Landau
Lifshitz
Fluid Mechanics
Third Revised English Edition
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Course of Theoretical Physics
Volume 6
L. D. Landau (Deceased) and E. M. Lifshitz
Institute of Physical Problems
USSR Academy of Sciences
gamon
CVII
COURSE OF THEORETICAL PHYSICS
Volume 6
FLUID MECHANICS
COURSE OF THEORETICAL PHYSICS
Vol. 1 Mechanics
Vol. 2 The Classical Theory of Fields
Vol. 3 Quantum Mechanics— NonRelativistic Theory
Vol. 4 Relativistic Quantum Theory
Vol. 5 Statistical Physics
Vol. 7 Tfceory o/ Elasticity
Vol. 8 Electrodynamics of Continuous Media
Vol. 9 Physical Kinetics
FLUID MECHANICS
by
L. D. LANDAU and E. M. LIFSHITZ
INSTITUTE OF PHYSICAL PROBLEMS, U.S.S.R. ACADEMY OF SCIENCES
Volume 6 of Course of Theoretical Physics
Translated from the Russian by
J. B. SYKES and W. H. REID
PERGAMON PRESS
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Copyright
©
1959
Pergamon Press Ltd.
First published in English 1959
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Third impression 1966
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Printed in Great Britain by J. W. Arrovismiih Ltd., Bristol
330/59
CONTENTS
Page
Preface to the English edition xi
Notation x ii
I. IDEAL FLUIDS
§1. The equation of continuity 1
§2. Euler's equation 2
§3. Hydrostatics 6
§4. The condition that convection is absent 8
§5. Bernoulli's equation 9
§6. The energy flux 10
§7. The momentum flux 12
§8 The conservation of circulation 14
§9. Potential flow 16
§10. Incompressible fluids 20
§11. The drag force in potential flow past a body 31
§12. Gravity waves 36
§13. Long gravity waves 42
§14. Waves in an incompressible fluid 44
II. VISCOUS FLUIDS
§15. The equations of motion of a viscous fluid 47
§16. Energy dissipation in an incompressible fluid 53
§17. Flow in a pipe 55
§18. Flow between rotating cylinders 60
§19. The law of similarity 61
§20. Stokes' formula 63
§21. The laminar wake 71
§22. The viscosity of suspensions 76
§23. Exact solutions of the equations of motion for a viscous fluid 79
§24. Oscillatory motion in a viscous fluid 88
§25. Damping of gravity waves 98
III. TURBULENCE
§26. Stability of steady flow 102
§27. The onset of turbulence 103
§28. Stability of flow between rotating cylinders 107
§29. Stability of flow in a pipe 1 1 1
vi Contents
Page
§30. Instability of tangential discontinuities 114
§31. Fully developed turbulence 116
§32. Local turbulence 120
§33. The velocity correlation 123
§34. The turbulent region and the phenomenon of separation 128
§35. The turbulent jet 130
§36. The turbulent wake 136
§37. Zhukovskii's theorem 137
§38. Isotropic turbulence 140
IV. BOUNDARY LAYERS
§39. The laminar boundary layer 145
§40. Flow near the line of separation 151
§41. Stability of flow in the laminar boundary layer 156
§42. The logarithmic velocity profile 159
§43. Turbulent flow in pipes 163
§44. The turbulent boundary layer 166
§45. The drag crisis 168
§46. Flow past streamlined bodies 172
§47. Induced drag 175
§48. The lift of a thin wing 179
V. THERMAL CONDUCTION IN FLUIDS
§49. The general equation of heat transfer 183
§50. Thermal conduction in an incompressible fluid 188
§51. Thermal conduction in an infinite medium 192
§52. Thermal conduction in a finite medium 196
§53. The similarity law for heat transfer 202
§54. Heat transfer in a boundary layer 205
§55. Heating of a body in a moving fluid 209
§56. Free convection 212
VI. DIFFUSION
§57. The equations of fluid dynamics for a mixture of fluids 219
§58. Coefficients of mass transfer and thermal diffusion 222
§59. Diffusion of particles suspended in a fluid 227
VII. SURFACE PHENOMENA
§60. Laplace's formula 230
§61. Capillary waves 237
§62. The effect of adsorbed films on the motion of a liquid 241
Contents vii
VIII. SOUND Page
§63. Sound waves 245
§64. The energy and momentum of sound waves 249
§65. Reflection and refraction of sound waves 253
§66. Geometrical acoustics 256
§67. Propagation of sound in a moving medium 259
§68. Characteristic vibrations 262
§69. Spherical waves 265
§70. Cylindrical waves 268
§71. The general solution of the wave equation 270
§72. The lateral wave 273
§73. The emission of sound 279
§74. The reciprocity principle 288
§75. Propagation of sound in a tube 291
§76. Scattering of sound 294
§77. Absorption of sound 298
§78. Second viscosity 304
IX. SHOCK WAVES
§79. Propagation of disturbances in a moving gas 310
§80. Steady flow of a gas 313
§81. Surfaces of discontinuity 317
§82. The shock adiabatic 319
§83. Weak shock waves 322
§84. The direction of variation of quantities in a shock wave 325
§85. Shock waves in a perfect gas 329
§86. Oblique shock waves 333
§87. The thickness of shock waves 337
§88. The isothermal discontinuity 342
§89. Weak discontinuities 344
X. ONEDIMENSIONAL GAS FLOW
§90. Flow of gas through a nozzle 347
§91. Flow of a viscous gas in a pipe 350
§92. Onedimensional similarity flow 353
§93. Discontinuities in the initial conditions 360
§94. Onedimensional travelling waves 366
§95. Formation of discontinuities in a sound wave 372
§96. Characteristics 373
§97. Riemann invariants 381
§98. Arbitrary onedimensional gas flow 386
§99. The propagation of strong shock waves 392
§100. Shallowwater theory 396
viii Contents
XL THE INTERSECTION OF SURFACES OF DISCONTINUITY
Page
§101. Rarefaction waves 399
§102. The intersection of shock waves 405
§103. The intersection of shock waves with a solid surface 410
§104. Supersonic flow round an angle 413
§105. Flow past a conical obstacle 418
XII. TWODIMENSIONAL GAS FLOW
§106. Potential flow of a gas 422
§107. Steady simple waves 425
§108. Chaplygin's equation: the general problem of steady two
dimensional gas flow 430
§109. Characteristics in steady twodimensional flow 433
§110. The EulerTricomi equation. Transonic flow 436
§111. Solutions of the EulerTricomi equation near nonsingular
points of the sonic surface 441
§1 12. Flow at the velocity of sound 446
§113. The intersection of discontinuities with the transition line 451
XIII. FLOW PAST FINITE BODIES
§114. The formation of shock waves in supersonic flow past bodies 457
§115. Supersonic flow past a pointed body 460
§116. Subsonic flow past a thin wing 464
§117. Supersonic flow past a wing 466
§118. The law of transonic similarity 469
§119. The law of hypersonic similarity 472
XIV. FLUID DYNAMICS OF COMBUSTION
§120. Slow combustion 474
§121. Detonation 480
§122. The propagation of a detonation wave 487
§123. The relation between the different modes of combustion 493
§124. Condensation discontinuities 496
XV. RELATIVISTIC FLUID DYNAMICS
§125. The energymomentum tensor 499
§126. The equations of relativistic fluid dynamics 500
§127. Relativistic equations for dissipative processes 505
Contents ix
XVI. DYNAMICS OF SUPERFLUIDS Page
§128. Principal properties of superfluids 507
§129. The thermomechanical effect 509
§130. The equations of superfluid dynamics 510
§131. The propagation of sound in a superfluid 517
XVII. FLUCTUATIONS IN FLUID DYNAMICS
§132. The general theory of fluctuations in fluid dynamics 523
§133. Fluctuations in an infinite medium 526
Index 530
PREFACE TO THE ENGLISH EDITION
The present book deals with fluid mechanics, i.e. the theory of the motion of
liquids and gases.
The nature of the book is largely determined by the fact that it describes
fluid mechanics as a branch of theoretical physics, and it is therefore markedly
different from other textbooks on the same subject. We have tried to develop
as fully as possible all matters of physical interest, and to do so in such a way
as to give the clearest possible picture of the phenomena and their interrela
tion. Accordingly, we discuss neither approximate methods of calculation in
fluid mechanics, nor empirical theories devoid of physical significance. On
the other hand, accounts are given of some topics not usually found in text
books on the subject: the theory of heat transfer and diffusion in fluids;
acoustics; the theory of combustion; the dynamics of superfluids; and
relativistic fluid dynamics.
In a field which has been so extensively studied as fluid mechanics it was
inevitable that important new results should have appeared during the several
years since the last Russian edition was published. Unfortunately, our
preoccupation with other matters has prevented us from including these
results in the English edition. We have merely added one further chapter,
on the general theory of fluctuations in fluid dynamics.
We should like to express our sincere thanks to Dr Sykes and Dr Reid
for their excellent translation of the book, and to Pergamon Press for their
ready agreement to our wishes in various matters relating to its publication.
Moscow L. D. Landau
E. M. Lifshitz
NOTATION
p density
p pressure
T temperature
s entropy per unit mass
e internal energy per unit mass
w = e +pjp heat function (enthalpy)
7 = c v jc v ratio of specific heats at constant pressure and constant
volume
rj dynamic viscosity
v = rjjp kinematic viscosity
k thermal conductivity
X = KJpCp thermometric conductivity
R Reynolds number
c velocity of sound
M ratio of fluid velocity to velocity of sound
CHAPTER I
IDEAL FLUIDS
§1. The equation of continuity
Fluid dynamics concerns itself with the study of the motion of fluids
(liquids and gases). Since the phenomena considered in fluid dynamics are
macroscopic, a fluid is regarded as a continuous medium. This means that
any small volume element in the fluid is always supposed so large that it still
contains a very great number of molecules. Accordingly, when we speak
of infinitely small elements of volume, we shall always mean those which are
"physically" infinitely small, i.e. very small compared with the volume of
the body under consideration, but large compared with the distances between
the molecules. The expressions fluid particle and point in a fluid are to be
understood in a similar sense. If, for example, we speak of the displacement
of some fluid particle, we mean not the displacement of an individual mole
cule, but that of a volume element containing many molecules, though still
regarded as a point.
The mathematical description of the state of a moving fluid is effected by
means of functions which give the distribution of the fluid velocity
v = v(x, y, z, t) and of any two thermodynamic quantities pertaining to
the fluid, for instance the pressure p(x, y, z, t) and the density p(x, y, z, t).
As is well known, all the thermodynamic quantities are determined by the
values of any two of them, together with the equation of state; hence, if
we are given five quantities, namely the three components of the velocity v,
the pressure p and the density p, the state of the moving fluid is completely
determined.
All these quantities are, in general, functions of the coordinates x, y, z
and of the time t. We emphasise that v(x, y, z y t) is the velocity of the
fluid at a given point (x, y, z) in space and at a given time t , i.e. it refers to
fixed points in space and not to fixed particles of the fluid ; in the course of
time, the latter move about in space. The same remarks apply to p and p.
We shall now derive the fundamental equations of fluid dynamics. Let
us begin with the equation which expresses the conservation of matter.
We consider some volume Vq of space. The mass of fluid in this volume
is / p d V, where p is the fluid density, and the integration is taken over the
volume Vo. The mass of fluid flowing in unit time through an element df
of the surface bounding this volume is pv • df ; the magnitude of the vector
df is equal to the area of the surface element, and its direction is along the
normal. By convention, we take df along the outward normal. Then p\ • df
is positive if the fluid is flowing out of the volume, and negative if the flow
1
2 Ideal Fluids §2
is into the volume. The total mass of fluid flowing out of the volume Vq
in unit time is therefore
<j> pv»df,
where the integration is taken over the whole of the closed surface surround
ing the volume in question.
Next, the decrease per unit time in the mass of fluid in the volume Vq
can be written
8t
Equating the two expressions, we have
d
~dt
!
8 .
pdV.
fpdF =  (jjpv.df. (1.1)
The surface integral can be transformed by Green's formula to a volume
integral:
(J) p v* df = div (pv) dV.
Thus
dp
J[^ + div(pv)]dF=0.
Since this equation must hold for any volume, the integrand must vanish,
i.e.
This is the equation of continuity. Expanding the expression div (pv), we
can also write (1.2) as
dp/dt + div (pv) = 0. (1.2)
ntinuity. Expanding the expression div (pv), we
dpjdt+p div v+vgradp = 0. (1.3)
The vector
J = pv (1.4)
is called the mass flux density. Its direction is that of the motion of the
fluid, while its magnitude equals the mass of fluid flowing in unit time
through unit area perpendicular to the velocity.
§2. Euler's equation
Let us consider some volume in the fluid. The total force acting on this
volume is equal to the integral
 jpdf
§2 Euler's equation 3
of the pressure, taken over the surface bounding the volume. Transforming
it to a volume integral, we have
 <j)/>df =  j gradpdV.
Hence we see that the fluid surrounding any volume element dV exerts
on that element a force dVgradp. In other words, we can say that a
force — gradp acts on unit volume of the fluid.
We can now write down the equation of motion of a volume element in
the fluid by equating the force gmdp to the product of the mass per unit
volume (/>) and the acceleration dvjdt:
p dvjdt = grad/>. (2.1)
The derivative dv/dt which appears here denotes not the rate of change
of the fluid velocity at a fixed point in space, but the rate of change of the
velocity of a given fluid particle as it moves about in space. This derivative
has to be expressed in terms of quantities referring to points fixed in space.
To do so, we notice that the change dv in the velocity of the given fluid
particle during the time dt is composed of two parts, namely the change
during d* in the velocity at a point fixed in space, and the difference between
the velocities (at the same instant) at two points dr apart, where dr is the
distance moved by the given fluid particle during the time dt. The first
part is (dv/dt)dt f where the derivative dv/dt is taken for constant x, y, z,
i.e. at the given point in space. The second part is
dv dv dv t%
dx— + dv— + dz— = (drgrad)v.
dx dy dz
Thus
dv = (dv/d*)d* + (drgrad)v,
or, dividing both sides by dt,
— = — +(vgrad)v. (2.2)
dt dt K
Substituting this in (2.1), we find
dv 1
— + (v«grad)v = gradp. (2.3)
dt p
This is the required equation of motion of the fluid; it was first obtained
by L. Euler in 1755. It is called Euler's equation and is one of the funda
mental equations of fluid dynamics.
If the fluid is in a gravitational field, an additional force />g, where g
is the acceleration due to gravity, acts on any unit volume. This force
4 Ideal Fluids §2
must be added to the righthand side of equation (2.1), so that equation (2.3)
takes the form
— + (v.grad)v =  5 L + g . (2 . 4 )
ot p
In deriving the equations of motion we have taken no account of processes
of energy dissipation, which may occur in a moving fluid in consequence of
internal friction (viscosity) in the fluid and heat exchange between different
parts of it. The whole of the discussion in this and subsequent sections of
this chapter therefore holds good only for motions of fluids in which thermal
conductivity and viscosity are unimportant; such fluids are said to be ideal.
The absence of heat exchange between different parts of the fluid (and
also, of course, between the fluid and bodies adjoining it) means that the
motion is adiabatic throughout the fluid. Thus the motion of an ideal
fluid must necessarily be supposed adiabatic.
In adiabatic motion the entropy of any particle of fluid remains constant
as that particle moves about in space. Denoting by s the entropy per unit
mass, we can express the condition for adiabatic motion as
dsjdt = 0, (2.5)
where the total derivative with respect to time denotes, as in (2.1), the rate
of change of entropy for a given fluid particle as it moves about. This
condition can also be written
ds/dt+vgrads *= 0. (2.6)
This is the general equation describing adiabatic motion of an ideal fluid.
Using (1.2), we can write it as an "equation of continuity" for entropy:
d(ps)jdt+div(psv) = 0. (2.7)
The product psv is the "entropy flux density".
It must be borne in mind that the adiabatic equation usually takes a much
simpler form. If, as usually happens, the entropy is constant throughout
the volume of the fluid at some initial instant, it retains everywhere the same
constant value at all times and for any subsequent motion of the fluid.
In this case we can write the adiabatic equation simply as
s = constant, (2.8)
and we shall usually do so in what follows. Such a motion is said to be
isentropic.
We may use the fact that the motion is isentropic to put the equation of
motion (2.3) in a somewhat different form. To do so, we employ the
familiar thermodynamic relation
dw = Tds+ Vdp,
where w is the heat function per unit mass of fluid (enthalpy), V = \jp
§2 Euler's equation 5
is the specific volume, and T is the temperature. Since s = constant, we
have simply
dzv = Vdp = dp/p,
and so (grad p)jp = grad w. Equation (2.3) can therefore be written in
the form
dv/d* + (vgrad)v = gradw. (2.9)
It is useful to notice one further form of Euler's equation, in which it in
volves only the velocity. Using a formula well known in vector analysis,
 grad v* = vxcurlv+(v»grad)v,
we can write (2.9) in the form
8v/dt+%gradv 2 YXC\irlv = gradw. (2.10)
If we take the curl of both sides of this equation, we obtain
8
— (curlv) = curl (v x curl v), (2.11)
8t
which involves only the velocity.
The equations of motion have to be supplemented by the boundary con
ditions that must be satisfied at the surfaces bounding the fluid. For an
ideal fluid, the boundary condition is simply that the fluid cannot penetrate
a solid surface. This means that the component of the fluid velocity normal
to the bounding surface must vanish if that surface is at rest:
v» = 0. (2.12)
In the general case of a moving surface, v n must be equal to the correspond
ing component of the velocity of the surface.
At a boundary between two immiscible fluids, the condition is that the
pressure and the velocity component normal to the surface of separation
must be the same for the two fluids, and each of these velocity components
must be equal to the corresponding component of the velocity of the
surface.
As has been said at the beginning of §1, the state of a moving fluid is
determined by five quantities : the three components of the velocity v and,
for example, the pressure p and the density p. Accordingly, a complete
system of equations of fluid dynamics should be five in number. For an
ideal fluid these are Euler's equations, the equation of continuity, and
the adiabatic equation.
PROBLEM
Write down the equations for onedimensional motion of an ideal fluid in terms of the
variables a, t, where a (called a Lagrangian variable^) is the * coordinate of a fluid particle
at some instant t = t .
t Although such variables are usually called Lagrangian, it should be mentioned that the equations
of motion in these coordinates were first obtained by Euler, at the same time as equations (2.3).
6 Ideal Fluids §3
Solution. In these variables the coordinate * of any fluid particle at any instant is re
garded as a function of t and its coordinate a at the initial instant: x = x(a, t). The condition
of conservation of mass during the motion of a fluid element (the equation of continuity)
is accordingly written p dx = p da, or
8x\
P'
\Sa/ L
where p {a) is a given initial density distribution. The velocity of a fluid particle is, by
definition, y = (dx[dt) a , and the derivative (8vjdt) a gives the rate of change of the velocity
of the particle during its motion. Euler's equation becomes
/8v\
and the adiabatic equation is
dv\ 1 /dp\
8t J a po \da)t
{ds/dt) a = 0.
§3. Hydrostatics
For a fluid at rest in a uniform gravitational field, Euler's equation (2.4)
takes the form
gradp = pg. (3.1)
This equation describes the mechanical equilibrium of the fluid. (If there
is no external force, the equation of equilibrium is simply gradp = 0,
i.e. p = constant; the pressure is the same at every point in the fluid.)
Equation (3.1) can be integrated immediately if the density of the fluid
may be supposed constant throughout its volume, i.e. if there is no signi
ficant compression of the fluid under the action of the external force. Taking
the sraxis vertically upward, we have
dp/dx = dpjdy = 0, Bpjdz = pg.
Hence
P — —pg z + constant.
If the fluid at rest has a free surface at height h, to which an external pressure
po, the same at every point, is applied, this surface must be the horizontal
plane z = h. From the condition^) = p for z = h, we find that the constant
is po + pgh y so that
p=po+ P g(hz). (3.2)
For large masses of liquid, and for a gas, the density p cannot in general
be supposed constant; this applies especially to gases (for example, the
atmosphere). Let us suppose that the fluid is not only in mechanical
equilibrium but also in thermal equilibrium. Then the temperature is the
§3 Hydrostatics 7
same at every point, and equation (3.1) may be integrated as follows. We
use the familiar thermodynamic relation
d<D = sdT+Vdp,
where <X> is the thermodynamic potential per unit mass. For constant tem
perature
d0> = Vdp = dpi p.
Hence we see that the expression (grad p)jp can be written in this case as
grad O, so that the equation of equilibrium (3.1) takes the form
grad$ = g.
For a constant vector g directed along the negative sraxis we have
g = grades).
Thus
grad(0+#sr) = 0,
whence we find that throughout the fluid
<!?+gz = constant; (3.3)
gz is the potential energy of unit mass of fluid in the gravitational field.
The condition (3.3) is known from statistical physics to be the condition
for thermodynamic equilibrium of a system in an external field.
We may mention here another simple consequence of equation (3.1).
If a fluid (such as the atmosphere) is in mechanical equilibrium in a gravi
tational field, the pressure in it can be a function only of the altitude z
(since, if the pressure were different at different points with the same alti
tude, motion would result). It then follows from (3.1) that the density
1 dp
gdz
is also a function of z only. The pressure and density together determine
the temperature, which is therefore again a function of z only. Thus, in
mechanical equilibrium in a gravitational field, the pressure, density and
temperature distributions depend only on the altitude. If, for example, the
temperature is different at different points with the same altitude, then
mechanical equilibrium is impossible.
Finally, let us derive the equation of equilibrium for a very large mass of
fluid, whose separate parts are held together by gravitational attraction —
a star. Let <£ be the Newtonian gravitational potential of the field due to
the fluid. It satisfies the differential equation
A<f> = 47rG/>, (3.5)
8 Ideal Fluids
§4
where G is the Newtonian constant of gravitation. The gravitational accelera
tion is grad <f>, and the force on a mass p is ~ P grad <f>. The condition
of equilibrium is therefore
gradp = pgrad<f>.
Dividing both sides by P , taking the divergence of both sides, and using
equation (3.5), we obtain
divgradpj = 4ttG p . (3.6)
It must be emphasised that the present discussion concerns only mechanical
equilibrium; equation (3.6) does not presuppose the existence of complete
thermal equilibrium.
If the body is not rotating, it will be spherical when in equilibrium,
and the density and pressure distributions will be spherically symmetrical.
Equation (3.6) in spherical coordinates then takes the form
1 d / r 2 dp \
§4. The condition that convection is absent
A fluid can be in mechanical equilibrium (i.e. exhibit no macroscopic
motion) without being in thermal equilibrium. Equation (3.1), the condi
tion for mechanical equilibrium, can be satisfied even if the temperature is
not constant throughout the fluid. However, the question then arises of
the stability of such an equilibrium. It is found that the equilibrium is
stable only when a certain condition is fulfilled. Otherwise, the equilibrium
is unstable, and this leads to the appearance in the fluid of currents which
tend to mix the fluid in such a way as to equalise the temperature. This
motion is called convection. Thus the condition for a mechanical equilibrium
to be stable is the condition that convection is absent. It can be derived as
follows.
Let us consider a fluid element at height z, having a specific volume
V(p, s), where p and s are the equilibrium pressure and entropy at height
z. Suppose that this fluid element undergoes an adiabatic upward displace
ment through a small interval g; its specific volume then becomes V(p\ s),
where p' is the pressure at height z + £. For the equilibrium to be stable, it
is necessary (though not in general sufficient) that the resulting force on
the element should tend to return it to its original position. This means
that the element must be heavier than the fluid which it "displaces" in its
new position. The specific volume of the latter is V(p\ s'), where s' is the
equilibrium entropy at height z+$. Thus we have the stability condition
V(p',s')V(p',s)>0.
§5 Bernoulli's equation 9
Expanding this difference in powers of s's = gdsjdz, we obtain
> 0. (4.1)
/ 8V \ ds
\ 8s Jp dz
dz
The formulae of thermodynamics give
8V\ T I 8V
p Cp \ 8 1 J p
where c v is the specific heat at constant pressure. Both c p and T are positive,
(v)fO
\ 8s 1 v Cp \ 81 ) p
it at
so that we can write (4.1) as
(^)^>0. (4.2)
\8Tlpdz
The majority of substances expand on heating, i.e. {8V]8T) V > 0. The
condition that convection is absent then becomes
ds/dz > 0, (4.3)
i.e. the entropy must increase with height.
From this we easily find the condition that must be satisfied by the
temperature gradient dTjdz. Expanding the derivative dsjdz, we have
ds _ / 8s \ dT /8s\ #_£p^_/j^\ ^ >0
dz = \~8f)p~dz~ + \8pl T dz~ T 7 dz \8T/pdz >
Finally, substituting from (3.4) dpjdz = g\V, we obtain
dT gT / 8V\
dz CpV \ 8T Jp
Convection can occur if the temperature falls with increasing height and the
magnitude of the temperature gradient exceeds (gTlc p V)(dV/8T)p.
If we consider the equilibrium of a column of a perfect gas, then
(T/V)(8V/8T)p = 1,
and the condition for stable equilibrium is simply
dTjdz > gjcp. (4.5)
§5. Bernoulli's equation
The equations of fluid dynamics are much simplified in the case of steady
flow. By steady flow we mean one in which the velocity is constant in time
at any point occupied by fluid. In other words, v is a function of the co
ordinates only, so that dvj8t = 0. Equation (2.10) then reduces to
grada 2 vxcurlv = gradw. (5.1)
We now introduce the concept of streamlines. These are lines such that
10 Ideal Fluids
§6
the tangent to a streamline at any point gives the direction of the velocity
at that point; they are determined by the following system of differential
equations :
dx dy dz
VX Vy Vz ^ ' '
In steady flow the streamlines do not vary with time, and coincide with the
paths of the fluid particles. In nonsteady flow this coincidence no longer
occurs: the tangents to the streamlines give the directions of the velocities
of fluid particles at various points in space at a given instant, whereas the
tangents to the paths give the directions of the velocities of given fluid
particles at various times.
We form the scalar product of equation (5.1) with the unit vector tangent
to the streamline at each point; this unit vector is denoted by 1. The pro
jection of the gradient on any direction is, as we know, the derivative in that
direction. Hence the projection of grad to is dtofdl. The vector vxcurl v
is perpendicular to v, and its projection on the direction of 1 is therefore
zero.
Thus we obtain from equation (5.1)
8
jfiv* + v>) = 0.
It follows from this that \v 2 +w is constant along a streamline:
%v 2 + w = constant. (5 # 3)
In general the constant takes different values for different streamlines.
Equation (5.3) is called Bernoulli's equation.
If the flow takes place in a gravitational field, the acceleration g due to
gravity must be added to the righthand side of equation (5.1). Let us take
the direction of gravity as the *axis, with z increasing upwards. Then the
cosine of the angle between the directions of g and 1 is equal to the derivative
— dsr/d/, so that the projection of g on 1 is
gdzfdl.
Accordingly, we now have
8
—(iv 2 + w+gz) = 0.
Thus Bernoulli's equation states that along a streamline
%v 2 +zo+gz = constant. (5.4)
§6. The energy flux
Let us choose some volume element fixed in space, and find how the
§6 The energy flux H
energy of the fluid contained in this volume element varies with time.
The energy of unit volume of fluid is
%pv 2 +pe,
where the first term is the kinetic energy and the second the internal energy,
€ being the internal energy per unit mass. The change in this energy is
given by the partial derivative
d
—$pv z + pe).
dt
To calculate this quantity, we write
d dp dv
— (W) = i^ 2 ^ + pV— ,
8f dt dt
or, using the equation of continuity (1.2) and the equation of motion (2.3),
o
— (ipv 2 ) = ^2div(pv)v.grad/>pv(v.grad)v.
dt "
In the last term we replace v • (v • grad)v by \v • grad v\ and grad/> by
p grad zvpT grad s (using the thermodynamic relation dw = Tds + (l/p)d/>),
obtaining
pi
—Qpv 2 ) = a 2 div(pv)pv.grad(^ 2 + «0+p:Zv.grad*.
dt
In order to transform the derivative d(pe)ldt, we use the thermodynamic
relation
de = TdspdV = Tds+(pl P 2 )dp.
Since €+pjp = e+pV is simply the heat function w per unit mass, we find
d[p € ) = edp + pde = zvdp+pTds,
and so
J^Z _ wJL + pT— = zo div (pv) pTv* grad s.
dt dt dt
Here we have also used the general adiabatic equation (2.6).
Combining the above results, we find the change in the energy to be
— Qpv^+pe) = (^ 2 +w)div(/>v)/>v.grad(^ 2 +w),
dt
or, finally,
p*
— (ipv 2 + pe) = div[pv(© 2 +w)]. (6.1)
dt
12 Ideal Fluids
§7
In order to see the meaning of this equation, let us integrate it over some
volume :
— J (ipv2+ P e)dV=  J div[pv$v2 + w)]dV,
or, converting the volume integral on the right into a surface integral,
— J (i P v2+pe)dV =  j>pv$v*+w).df. ^
The lefthand side is the rate of change of the energy of the fluid in some
given volume. The righthand side is therefore the amount of energy
flowing out of this volume in unit time. Hence we see that the expression
pv{\v* + w) ( 6<3 )
may be called the energy flux density vector. Its magnitude is the amount of
energy passing in unit time through unit area perpendicular to the direction
of the velocity.
The expression (6.3) shows that any unit mass of fluid carries with it during
its motion an amount of energy w + %v\ The fact that the heat function w
appears here, and not the internal energy e, has a simple physical signifi
cance. Putting w = e+plp, we can write the flux of energy through a closed
surface in the form
 j>pv(±v 2 + e)df jpvdf.
The first term is the energy (kinetic and internal) transported through the
surface in unit time by the mass of fluid. The second term is the work done
by pressure forces on the fluid within the surface.
§7. The momentum flux
We shall now give a similar series of arguments for the momentum of the
fluid. The momentum of unit volume is pv. Let us determine its rate of
change, 8(pv)/dt. We shall use tensor notation.f We have
8 dvt 8p
(^) = /) __ + _^
a J J he La ? n S "f xes *' *'  take the v f Iues 1. 2, 3, corresponding to the components of vectors
and tensors along the axes x, y, z respectively. We shall write sums of the type A • B = A.B,+ AB +
a ~ ~f ^ * n the . form AfBt simply, omitting the summation sign. We shall use a similar pro
cedure in all products involving vectors or tensors: summation over the values 1, 2, 3 is always under
stood when a Latin suffix appears twice in any term. Such suffixes are sometimes called dummy
suffixes In working with dummy suffixes it should be remembered that any pair of such suffixes may
be replaced by any other like letters, since the notation used for suffixes that take all possible values
obviously does not affect the value of the sum.
§7 The momentum flux 13
Using the equation of continuity (1.2) (with div(pv) written in the form
d{pv k )jdx k )
dp d(pv k )
dt dx k '
and Euler's equation (2.3) in the form
dv{ dvi 1 dp
= — V k ,
dt dxjc p dxt
we obtain
d dvt dp d(pv k )
(pVi) = — pv k Vi—
dV dx k dxt dx k
=  —  —ipViV k ).
dxi dx k
We write the first term on the right in the formf
dp dp
—— = oi k  — ,
dxi dx k
and finally obtain
d dU ik
where the tensor Ui k is defined as
Ilta = p8 ik +pviv k . (7.2)
This tensor is clearly symmetrical.
To see the meaning of the tensor 11^, we integrate equation (7.1) over some
volume :
dt] H J dx k
The integral on the right is transformed into a surface integral by Green's
formula :%
j (pv t dV= j>U ik df k . (7.3)
t 8,* denotes the unit tensor, i.e. the tensor with components which are unity for i ; = k and zero
for »' 4= k. It is evident that h i]c A k = A it where A t is any vector. Similarly, if Am is a tensor of rank
two, we have the relations BaAia = An, BaAa = A iit and so on.
J The rule for transforming an integral over a closed surface into one over the volume bounded
by that surface can be formulated as follows: the surface element d/ f must be replaced by the operator
dV • 8/dxi, which is to be applied to the whole of the integrand.
14 Ideal Fluids
§8
The lefthand side is the rate of change of the *th component of the
momentum contained in the volume considered. The surface integral on
the right is therefore the amount of momentum flowing out through the
bounding surface in unit time. Consequently, U ik df k is the rth component
of the momentum flowing through the surface element d/. If we write d/*
in the form n k d/, where d/is the area of the surface element, and n is a unit
vector along the outward normal, we find that Yl ik n k is the flux of the tth
component of momentum through unit surface area. We may notice that,
according to (7.2), Il ik n k = pn t + P ViV k n k . This expression can be written
in vector form
pn+ P v(vn). (7.4)
Thus Tl ik is the jth component of the amount of momentum flowing in
unit time through unit area perpendicular to the a^axis. The tensor U ik
is called the momentum flux density tensor. The energy flux is determined by
a vector, energy being a scalar; the momentum flux, however, is determined
by a tensor of rank two, the momentum itself being a vector.
The vector (7.4) gives the momentum flux in the direction of n, i.e.
through a surface perpendicular to n. In particular, taking the unit vector
n to be directed parallel to the fluid velocity, we find that only the longitu
dinal component of momentum is transported in this direction, and its
flux density is p + pv 2 . In a direction perpendicular to the velocity, only the
transverse component (relative to v) of momentum is transported, its flux
density being just p.
§8. The conservation of circulation
The integral
T = jvdl
taken along some closed contour, is called the velocity circulation round that
contour.
Let us consider a closed contour drawn in the fluid at some instant.
We suppose it to be a "fluid contour", i.e. composed of the fluid particles
that lie on it. In the course of time these particles move about, and the
contour moves with them. Let us investigate what happens to the velocity
circulation. In other words, let us calculate the time derivative
Tt§
vdl.
We have written here the total derivative with respect to time, since we are
seeking the change in the circulation round a "fluid contour" as it moves
about, and not round a contour fixed in space.
To avoid confusion, we shall temporarily denote differentiation with respect
§8 The conservation of circulation 15
to the coordinates by the symbol S, retaining the symbol d for differentia
tion with respect to time. Next, we notice that an element dl of the length
of the contour can be written as the difference Sr between the radius vectors
r of the points at the ends of the element. Thus we write the velocity cir
culation as § v • or. In differentiating this integral with respect to time, it
must be borne in mind that not only the velocity but also the contour itself
(i.e. its shape) changes. Hence, on taking the time differentiation under the
integral sign, we must differentiate not only v but also Sr:
d r f dv r dSr
— * v«Sr = (b — «Sr + <b v— — .
dtj 7 dt 7 dt
Since the velocity v is just the time derivative of the radius vector r,
we have
v .^ = v.S— = vSv = S(i* 2 ).
dt dt ^
The integral of a total differential along a closed contour, however, is zero.
The second integral therefore vanishes, leaving
d r w r dv
— d) v«or = d) ——or.
dt 7 7 dt
It now remains to substitute for the acceleration dv[dt its expression
from (2.9):
dv/dt = — gradw.
Using Stokes' formula, we then have
since curl grad w = 0. Thus, going back to our previous notation, we
findf
or
<j>vdl = constant. (81)
We have therefore reached the conclusion that, in an ideal fluid, the velocity
circulation round a closed "fluid" contour is constant in time {Kelvin's
theorem or the law of conservation of circulation).
It should be emphasised that this result has been obtained by using Euler's
equation in the form (2.9), and therefore involves the assumption that the
t This result remains valid in a uniform gravitational field, since in that case curl g = 0.
16 Ideal Fluids §9
flow is isentropic. The theorem does not hold for flows which are not
isentropic.f
§9. Potential flow
From the law of conservation of circulation we can derive an important
result. Let us at first suppose that the flow is steady, and consider a stream
line of which we know that to = curl v (the vorticity) is zero at some
point. We draw an arbitrary infinitely small closed contour to encircle the
streamline at that point. By Stokes' theorem, the velocity circulation round
any infinitely small contour is equal to curl v . df, where df is the element of
area enclosed by the contour. Since the contour at present under considera
tion is situated at a point where to = 0, the velocity circulation round it is
zero. In the course of time, this contour moves with the fluid, but always
remains infinitely small and always encircles the same streamline. Since
the velocity circulation must remain constant, i.e. zero, it follows that o
must be zero at every point on the streamline.
Thus we reach the conclusion that, if at any point on a streamline to = 0,
the same is true at all other points on that streamline. If the flow is not
steady, the same result holds, except that instead of a streamline we must
consider the path described in the course of time by some particular fluid
particle; J we recall that in nonsteady flow these paths do not in general
coincide with the streamlines.
At first sight it might seem possible to base on this result the following
argument. Let us consider steady flow past some body. Let the incident
flow be uniform at infinity; its velocity v is a constant, so that co s on all
streamlines. Hence we conclude that to is zero along the whole of every
streamline, i.e. in all space.
A flow for which to = in all space is called a potential flow or irrotational
flow, as opposed to rotational flow, in which the vorticity is not everywhere
zero. Thus we should conclude that steady flow past any body, with a
uniform incident flow at infinity, must be potential flow.
Similarly, from the law of conservation of circulation, we might argue
as follows. Let us suppose that at some instant we have potential flow
throughout the volume of the fluid. Then the velocity circulation round any
closed contour in the fluid is zero.ff By Kelvin's theorem, we could then
conclude that this will hold at any future instant, i.e. we should find that, if
t Mathematically, it is necessary that there should be a onetoone relation between/) and p (which
tor isentropic flow is s(p, p)  constant); then (1/p) grad/> can be written as the gradient of some
function, a result which is needed in deriving Kelvin's theorem.
J To avoid misunderstanding, we may mention here that this result has no meaning in turbulent
flow (cf. Chapter III). We may also remark that a nonzero vorticity may occur on a streamline after
the passage of a shock wave. We shall see that this is because the flow is no longer isentropic, and the
law of conservation of circulation cannot then be derived (§106).
tt Here we suppose for simplicity that the fluid occupies a simplyconnected region of space The
same final result would be obtained for a multiplyconnected region, but restrictions on the choice of
contours would have to be made in the derivation.
§9
Potential flow
17
there is potential flow at some instant, then there is potential flow at all
subsequent instants (in particular, any flow for which the fluid is initially
at rest must be a potential flow). This is in accordance with the fact that,
if to = 0, equation (2.11) is satisfied identically.
In fact, however, all these conclusions are of only very limited validity.
The reason is that the proof given above that to = all along a streamline
is, strictly speaking, invalid for a line which lies in the surface of a solid
body past which the flow takes place, since the presence of this surface makes
it impossible to draw a closed contour in the fluid encircling such a stream
line. The equations of motion of an ideal fluid therefore admit solutions for
which separation occurs at the surface of the body: the streamlines, having
followed the surface for some distance, become separated from it at some
point and continue into the fluid. The resulting flow pattern is characterised
by the presence of a "surface of tangential discontinuity" proceeding from
the body; on this surface the fluid velocity, which is everywhere tangential
to the surface, has a discontinuity. In other words, at this surface one layer
of fluid "slides" on another. Fig. 1 shows a surface of discontinuity which
separates moving fluid from a region of stationary fluid behind the body.
Fig. 1
From a mathematical point of view, the discontinuity in the tangential velocity
component corresponds to a surface on which the vorticity is nonzero.
When such discontinuous flows are included, the solution of the equations
of motion for an ideal fluid is not unique : besides continuous flow, they admit
also an infinite number of solutions possessing surfaces of tangential dis
continuity starting from any prescribed line on the surface of the body
past which the flow takes place. It should be emphasised, however, that
none of these discontinuous solutions is physically significant, since tangen
tial discontinuities are wholly unstable, and therefore the flow would in fact
become turbulent (see Chapter III).
The actual physical problem of flow past a given body has, of course, a
unique solution. The reason is that ideal fluids do not really exist; any
actual fluid has a certain viscosity, however small. This viscosity may have
practically no effect on the motion of most of the fluid, but, no matter how
small it is, it will be important in a thin layer of fluid adjoining the body.
18 Ideal Fluids §9
The properties of the flow in this boundary layer decide the choice of one out
of the infinity of solutions of the equations of motion for an ideal fluid.
It is found that, in the general case of flow past bodies of arbitrary form,
solutions with separation must be rejected; separation, if it occurred, would
result in turbulence.
In spite of what we have said above, the study of the solutions of the
equations of motion for continuous steady potential flow past bodies is in
some cases meaningful. Although, in the general case of flow past bodies of
arbitrary form, the actual flow pattern bears almost no relation to the pattern
of potential flow, for bodies of certain special ("streamlined"— §46) shapes
the flow may differ very little from potential flow; more precisely, it will be
potential flow except in a thin layer of fluid at the surface of the body and in
a relatively narrow "wake" behind the body.
Another important case of potential flow occurs for small oscillations of
a body immersed in fluid. It is easy to show that, if the amplitude a of the
oscillations is small compared with the linear dimension / of the body
(a <^ /), the flow past the body will be potential flow. To show this, we esti
mate the order of magnitude of the various terms in Euler's equation
dv/dt+(vgrad)v = — gradw.
The velocity v changes markedly (by an amount of the same order as the
velocity u of the oscillating body) over a distance of the order of the dimen
sion / of the body. Hence the derivatives of v with respect to the coordinates
are of the order of u\l. The order of magnitude of v itself (at fairly small
distances from the body) is determined by the magnitude of u. Thus we
have (v • grad)v ~ u 2 /l. The derivative dvjdt is of the order of <ou, where
o> is the frequency of the oscillations. Since w ~ uja, we have dvjdt ~ u 2 /a.
It now follows from the inequality a <^l that the term (v • grad)v is small
compared with dvjdt and can be neglected, so that the equation of motion
of the fluid becomes dvjdt = grad zo. Taking the curl of both sides, we
obtain d(curl v)jdt = 0, whence curl v = constant. In oscillatory motion,
however, the time average of the velocity is zero, and therefore curl v
= constant implies that curl v = 0. Thus the motion of a fluid executing
small oscillations is potential flow to a first approximation.
We shall now obtain some general properties of potential flow. We first
recall that the derivation of the law of conservation of circulation, and there
fore all its consequences, were based on the assumption that the flow is
isentropic. If the flow is not isentropic, the law does not hold, and therefore,
even if we have potential flow at some instant, the vorticity will in general
be nonzero at subsequent instants. Thus only isentropic flow can in fact
be potential flow.
According to Stokes' theorem,
<pv«dl = <j>curlv»df,
where the integral on the right is taken over a surface bounded by the contour
§9 Potential flow 19
in question. Hence we see that, in potential flow, the velocity circulation
round any closed contour is zero:
£vdl = 0. (9.1)
It follows from this that, in particular, closed streamlines cannot exist in
potential flow.f For, since the direction of a streamline is at every point
the direction of the velocity, the circulation along such a line can never be
zero.
In rotational motion the velocity circulation is not in general zero. In
this case there may be closed streamlines, but it must be emphasised that the
presence of closed streamlines is not a necessary property of rotational
motion.
Like any vector field having zero curl, the velocity in potential flow can
be expressed as the gradient of some scalar. This scalar is called the velocity
potential', we shall denote it by (f>:
v = grad<£. (9.2)
Writing Euler's equation in the form (2.10)
0v/d*+grada 2 vxcurlv = gradw
and substituting v = grad <f>, we have
grad I — + \v 2 +w I = 0,
whence
d<f>!dt+^v 2 +w =f(t), (9.3)
where f(t) is an arbitrary function of time. This equation is a first integral
of the equations of potential flow. The function /(*) in equation (9.3) can
be put equal to zero without loss of generality. For, since the velocity is
the space derivative of <j>, we can add to <f> any function of the time; replacing
by <f> + $f(t)dt, we obtain zero on the righthand side of (9.3).
For steady flow we have (taking the potential cf> to be independent of time)
dcf>jdt = 0, f{t) = constant, and (9.3) becomes Bernoulli's equation:
\v 2 +w = constant. (9.4)
It must be emphasised here that there is an important difference between the
Bernoulli's equation for potential flow and that for other flows. In the
general case, the "constant" on the righthand side is a constant along any
given streamline, but is different for different streamlines. In potential flow,
t This result, like (9.1), may not be valid for motion in a multiplyconnected region of space.
In potential flow in such a region, the velocity circulation may be nonzero if the closed contour
round which it is taken cannot be contracted to a point without crossing the boundaries of the region.
20 Ideal Fluids
§10
however, it is constant throughout the fluid. This enhances the importance
of Bernoulli's equation in the study of potential flow.
§10. Incompressible fluids
In a great many cases of the flow of liquids (and also of gases), their
density may be supposed invariable, i.e. constant throughout the volume of
the fluid and throughout its motion. In other words, there is no noticeable
compression or expansion of the fluid in such cases. We then speak of
incompressible flow.
The general equations of fluid dynamics are much simplified for an
incompressible fluid. Euler's equation, it is true, is unchanged if we put
p = constant, except that p can be taken under the gradient operator in
equation (2.4) :
— + (vgrad)v = grad/  j +g.
(10.1)
The equation of continuity, on the other hand, takes for constant p the
simple form
div v = 0. (10.2)
Since the density is no longer an unknown function as it was in the general
case, the fundamental system of equations in fluid dynamics for an incom
pressible fluid can be taken to be equations involving the velocity only
These may be the equation of continuity (10.2) and equation (2.11):
8
— (curl v) = curl <Vx curl v). (10.3)
Bernoulli's equation can be written in a simpler form for an incompressible
fluid. Equation (10.1) differs from the general Euler's equation (2.9) in that
it has grad {pip) in place of grad w. Hence we can write down Bernoulli's
equation immediately by simply replacing the heat function in (5.4) by pjp:
%v 2 +p/p+gz = constant. (10.4)
For an incompressible fluid, we can also write pjp in place of w in the
expression (6.3) for the energy flux, which then becomes
{v + >\
pv\&*+y (io.5)
For we have, from a wellknown thermodynamic relation, the expression
de = TdspdV for the change in internal energy; for s = constant and
V ~ \Jp = constant, de = 0, i.e. e = constant. Since constant terms in
the energy do not matter, we can omit e in zv = e+pjp.
§10
Incompressible fluids
21
The equations are particularly simple for potential flow of an incom
pressible fluid. Equation (10.3) is satisfied identically if curl v = 0. Equa
tion (10.2), with the substitution v = grad <f>, becomes
A<£ = 0,
(10.6)
i.e. Laplace's equation^ for the potential <f>. This equation must be supple
mented by boundary conditions at the surfaces where the fluid meets solid
bodies. At fixed solid surfaces, the fluid velocity component v n normal to
the surface must be zero, whilst for moving surfaces it must be equal to the
normal component of the velocity of the surface (a given function of time).
The velocity v n , however, is equal to the normal derivative of the potential
<f> : v n = 8<f>ldn. Thus the general boundary conditions are that dj>\dn is
a given function of coordinates and time at the boundaries.
For potential flow, the velocity is related to the pressure by equation (9.3).
In an incompressible fluid, we can replace zo in this equation by pjp :
a+iat+w+p/p = /(')• (10.7)
We may notice here the following important property of potential flow of
an incompressible fluid. Suppose that some solid body is moving through
the fluid. If the result is potential flow, it depends at any instant only on
the velocity of the moving body at that instant, and not, for example, on its
acceleration. For equation (10.6) does not explicitly contain the time, which
enters the solution only through the boundary conditions, and these contain
only the velocity of the moving body.
Fig. 2
From Bernoulli's equation, %v 2 +plp = constant, we see that, in steady
flow of an incompressible fluid (not in a gravitational field), the greatest
pressure occurs at points where the velocity is zero. Such a point usually
occurs on the surface of a body past which the fluid is moving (at the point
O in Fig. 2), and is called a stagnation point. If u is the velocity of the
t The velocity potential was first introduced by Euler, who obtained an equation of the form
(10.6) for it; this form later became known as Laplace's equation.
22 Ideal Fluids §10
incident current (i.e. the fluid velocity at infinity), and po the pressure at
infinity, the pressure at the stagnation point is
pmzx = po + lpuK (10.8)
If the velocity distribution in a moving fluid depends on only two co
ordinates (x and y, say), and the velocity is everywhere parallel to the
ryplane, the flow is said to be twodimensional or plane flow. To solve
problems of twodimensional flow of an incompressible fluid, it is sometimes
convenient to express the velocity in terms of what is called the stream
function. From the equation of continuity divv = dvxjdx+dvy/dy = we
see that the velocity components can be written as the derivatives
v x = difjjdy, v y = di/jjdx (10.9)
of some function ip(x, y), called the stream function. The equation of con
tinuity is then satisfied automatically. The equation that must be satisfied
by the stream function is obtained by substituting (10.9) in equation (10.3).
We then obtain
d 8ib 8 8ib 8
* A * a 7a7 A * + ^ A * = ° (1<U0)
If we know the stream function we can immediately determine the form of
the streamlines for steady flow. For the differential equation of the stream
lines (in twodimensional flow) is dxjv x = dyjv y or v y dx — v x dy = ;
it expresses the fact that the direction of the tangent to a streamline is the
direction of the velocity. Substituting (10.9), we have
difj 8i[i
— dx H dy = d^r = 0,
dx dy
whence j/t == constant. Thus the streamlines are the family of curves obtained
by putting the stream function \p{x, y) equal to an arbitrary constant.
If we draw a curve between two points A and B in the aryplane, the mass
flux Q across this curve is given by the difference in the values of the stream
function at these two points, regardless of the shape of the curve. For, if
v n is the component of the velocity normal to the curve at any point, we have
B B B
Q = p fv n dl = p j> (v y dx+v x dy) = p \ dip,
A A A
or
Q = MbIa). (10.11)
There are powerful methods of solving problems of twodimensional poten
tial flow of an incompressible fluid past bodies of various profiles, involving
§10 Incompressible fluids 23
the application of the theory of functions of a complex variable.f The basis
of these methods is as follows. The potential and the stream function are
related to the velocity components by
v x = dcf>/dx = difi/dy, v y = 8<f>jdy =  dip/dx.
These relations between the derivatives of <j> and ijj, however, are the same,
mathematically, as the wellknown CauchyRiemann conditions for a complex
expression
w = <f> + irfi (10.12)
to be an analytic function of the complex argument z = x+iy. This means
that the function w(z) has at every point a welldefined derivative
dw d<f> dift
— — = \ i — = v x —Wy. (10.13)
dz 8x 8x y v '
The function to is called the complex potential, and dwfdz the complex velocity.
The modulus and argument of the latter give the magnitude v of the velocity
and the angle 6 between the direction of the velocity and that of the #axis :
dzvjdz = ve*°. (10.14)
At a solid surface past which the flow takes place, the velocity must be
along the tangent. That is, the profile contour of the surface must be a
streamline, i.e. ^ = constant along it; the constant may be taken as zero,
and then the problem of flow past a given contour reduces to the deter
mination of an analytic function tv(z) which takes real values on the contour.
The statement of the problem is more involved when the fluid has a free
surface ; an example is found in Problem 9.
The integral of an analytic function round any closed contour C is well
known to be equal to 2rri times the sum of the residues of the function at its
simple poles inside C; hence
<J> w'dz = 2rri ^ Afo
k
where Ak are the residues of the complex velocity. We also have
<j> w' dz = <x> (v x — iv y )(dx + idy)
= <j> (v x dx + v y dy) + ij> (v x dy — v y dx).
f A more detailed account of these methods and their various applications is given by N. E. Kochin,
I. A. Kibel' and N. V. Roze, Theoretical Hydromechanics (Teoreticheskaya gidromekhanika), Part 1,
4th ed., Moscow 1948; L. I. Sedov, Twodimensional Problems of Hydrodynamics and Aerodynamics
(Ploskie zadachi gidrodinamxki i a'erodinamiki), Moscow 1950.
24 Ideal Fluids §10
The real part of this expression is just the velocity circulation V round
the contour C. The imaginary part, multiplied by p, is the mass flux across
C; if there are no sources of fluid within the contour, this flux is zero and
we then have simply
r = 2in^A k ; (10.15)
all the residues Ajc are in this case purely imaginary.
Finally, let us consider the conditions under which the fluid may be
regarded as incompressible. When the pressure changes adiabatically by
Ap, the density changes by Ap = {dpjdp) 8 Ap. According to Bernoulli's
equation, however, Ap is of the order of pv 2 in steady flow. Thus Ap ~
{dpjdp) s pv 2 . We shall show in §63 that the derivative (8pjdp) s is the square
of the velocity c of sound in the fluid, so that Ap ~ pv 2 jc 2 . The fluid may be
regarded as incompressible if Ap/p <^ 1. We see that a necessary condition
for this is that the fluid velocity should be small compared with that of
sound :
v <£ c. (10.16)
However, this condition is sufficient only in steady flow. In nonsteady
flow, a further condition must be fulfilled. Let t and / be a time and a length
of the order of the times and distances over which the fluid velocity undergoes
significant changes. If the terms dvjdt and (l//>) gradp in Euler's equation
are comparable, we find, in order of magnitude, vjr ~ Apjlp or Ap ~ Ipvjr,
and the corresponding change in p is Ap ~ Ipvjrc 2 . Now comparing the terms
dp/dt and p div v in the equation of continuity, we find that the derivative
dpjdt may be neglected (i.e. we may suppose p constant) if Apjr <^ pvjl,
or
t > Ijc. (10.17)
If the conditions (10.16) and (10.17) are both fulfilled, the fluid may be
regarded as incompressible. The condition (10.17) has an obvious meaning:
the time Ijc taken by a sound signal to traverse the distance / must be small
compared with the time t during which the flow changes appreciably, so
that the propagation of interactions in the fluid may be regarded as instan
taneous.
PROBLEMS
Problem 1. Determine the shape of the surface of an incompressible fluid subject to a
gravitational field, contained in a cylindrical vessel which rotates about its (vertical) axis with
a constant angular velocity fi.
Solution. Let us take the axis of the cylinder as the araxis. Then vx = — yCl, v y = *Q,
§10 Incompressible fluids 25
vz = 0. The equation of continuity is satisfied identically, and Euler's equation (10.1)
gives
p ox p ay p oz
The general integral of these equations is
p/p = %Q. 2 (x 2 +y 2 )—gz+ constant.
At the free surface p = constant, so that the surface is a paraboloid:
z = & 2 (x 2 +y 2 )/g t
the origin being taken at the lowest point of the surface.
Problem 2. A sphere, of radius R, moves with velocity u in an incompressible ideal fluid.
Determine the potential flow of the fluid past the sphere.
Solution. The fluid velocity must vanish at infinity. The solutions of Laplace's equation
A <f> = which vanish at infinity are well known to be ljr and the derivatives, of various orders,
of 1/r with respect to the coordinates (the origin is taken at the centre of the sphere). On
account of the complete symmetry of the sphere, only one constant vector, the velocity u,
can appear in the solution, and, on account of the linearity of both Laplace's equation and
the boundary condition, <f> must involve u linearly. The only scalar which can be formed
from u and the derivatives of 1/r is the scalar product u • grad(l/r). We therefore seek ^
in the form
<f> = A.grad(l/r) = (A.n)/r2,
where n is a unit vector in the direction of r. The constant A is determined from the condition
that the normal components of the velocities v and u must be equal at the surface at the
sphere, i.e. vn = u*n for r = R. This condition gives A = iuR 3 , so that
The pressure distribution is given by equation (10.7) :
P = po^pv 2 p8<f>Jdt,
where p is the pressure at infinity. To calculate the derivative 8<f>[8t, we must bear in mind
that the origin (which we have taken at the centre of the sphere) moves with velocity u.
Hence
d<f>/dt = (d(f>/du)uwgrad(f).
The pressure distribution over the surface of the sphere is given by the formula
P = po+$pu 2 (9 cos 2 d5)+ipRn'du/dt,
where 6 is the angle between n and u.
Problem 3. The same as Problem 2, but for an infinite cylinder moving perpendicular to
its axis.f
t The solution of the more general problems of potential flow past an ellipsoid and an elliptical
cylinder may be found in: N. E. Kochin, I. A. Kibel' and N. V. Roze, Theoretical Hydromechanics
(Teoreticheskaya gidromekhanika), Part 1, 4th ed., pp. 265 and 355, Moscow 1948; H. Lamb, Hydro
dynamics, 6th ed., §§103116, Cambridge 1932.
26 Ideal Fluids §10
Solution. The flow is independent of the axial coordinate, so that we have to solve
Laplace's equation in two dimensions. The solutions which vanish at infinity are the first
and higher derivatives of log r with respect to the coordinates, where r is the radius vector
perpendicular to the axis of the cylinder. We seek a solution in the form
<£ = A«gradlogr = A»n/r,
and from the boundary conditions we obtain A = — B?u, so that
R 2 R 2
9= u«n, v = ~[2n(u.n)u].
The pressure at the surface of the cylinder is given by the formula
P = A>+pw 2 (4 cos 2 d3)+ p RnduJdt.
Problem 4. Determine the potential flow of an incompressible ideal fluid in an ellipsoidal
vessel rotating about a principal axis with angular velocity Q, and determine the total angular
momentum of the fluid.
Solution. We take Cartesian coordinates x, y, z along the axes of the ellipsoid at a given
instant, the zaxis being the axis of rotation. The velocity of points in the vessel is
u = SI x r,
so that the boundary condition v n = d<$>\dn = «„ is
d$\dn = Q(xn y —yn x ),
or, using the equation of the ellipsoid x 2 /a 2 +y 2 Jb 2 +z 2 /c 2 = 1,
x d(f>
a 2 8x b 2 dy
The solution of Laplace's equation which satisfies this boundary condition is
a 2 b 2
* = Q a^b 2Xy ' (1)
The angular momentum of the fluid in the vessel is
M = p J (xv y yv x )dV.
Integrating over the volume V of the ellipsoid, we have
QpV (cfib 2 ) 2
y d 9 z 8 9 / 1 1 \
b 2 dy c 2 8z J \ b 2 a 2 J
M =
a 2 +b 2
Formula (1) gives the absolute motion of the fluid relative to the instantaneous position
of the axes x, y, z which are fixed to the rotating vessel. The motion relative to the vessel
(i.e. relative to a rotating system of coordinates *, y, z) is found by subtracting the velocity
SiXr from the absolute velocity; denoting the relative velocity of the fluid by v', we have
d 9 ^ 2Q.a 2 , 2Q&2
v x = _ + y a = v, v ' =  v v' z = o.
dx a 2 + b 2 a 2 + b 2
The paths of the relative motion are found by integrating the equations x = v' x , y = v' y ,
and are the ellipses x s /a z +y 2 /b 2 = constant, which are similar to the boundary ellipse.
§10 Incompressible fluids 27
Problem 5. Determine the flow near a stagnation point (Fig. 2).
Solution. A small part of the surface of the body near the stagnation point may be
regarded as plane. Let us take it as the aryplane. Expanding <f> for *, y, z small, we have
as far as the secondorder terms
<j> = ax+by + cz+Ax 2 +By 2 + Cz 2 +Dxy + Eyz+Fzx;
a constant term in <f> is immaterial. The constant coefficients are determined so that <f> satisfies
the equation A^ = and the boundary conditions v t = d<f>/dz = for z — and all x, y,
d<f>Jdx = d<f>/dy — 0fotx=y = z = (the stagnation point). This gives a = b = c = 0;
C = —A —B, E = F = 0. The term Dxy can always be removed by an appropriate rotation
of the * and y axes. We then have
<f> = Ax 2 + By 2 (A + B)z 2 . (1)
If the flow is axially symmetrical about the saxis (symmetrical flow past a solid of revo
lution), we must have A = B, so that
<f> = A{x 2 +y 2 2z 2 ).
The velocity components are v x = 2Ax, v v = 2Ay, v z =■ —\Az. The streamlines are given
by equations (5.2), from which we find x 2 z = c lt y 2 z = c it i.e. the streamlines are cubical
hyperbolae.
If the flow is uniform in the ydirection (e.g. flow in the #direction past a cylinder with
its axis in the ydirection), we must have B = in (1), so that
<f> = A{x 2 z*).
The streamlines are the hyperbolae xz = constant.
Problem 6. Determine the potential flow near an angle formed by two intersecting
planes.
Solution. Let us take polar coordinates r, 6 in the crosssectional plane (perpendicular
to the line of intersection), with the origin at the vertex of the angle ; 6 is measured from one
of the arms of the angle. Let the angle be a radians ; for a < it the flow takes place within
the angle, for a > v outside it. The boundary condition that the normal velocity component
vanishes means that 8<f>l8B = for 6 = and = a. The solution of Laplace's equation
satisfying these conditions can be written!
<f> = Ar n cos nd, n — ir/a,
so that
v r = nAr 11  1 cos nd, v d = — nAr n smnd.
For « < 1 (flow outside an angle; Fig. 3), v r becomes infinite as l/r 1_n at the origin. For
n > 1 (flow inside an angle; Fig. 4), v T becomes zero for r = 0.
The stream function, which gives the form of the streamlines, is tjt = Ar n sin nd. The
expressions obtained for <j> and ^ are the real and imaginary parts of the complex potential
to = Az n .
Problem 7. A spherical hole of radius a is suddenly formed in an incompressible fluid
filling all space. Determine the time taken for the hole to be filled with fluid (Rayleigh
1917).
Solution. The flow after the formation of the hole will be spherically symmetrical, the
t We take the solution which involves the lowest positive power of r, since r is small.
28
Ideal Fluids
§10
velocity at every point being directed to the centre of the hole. For the radial velocity
»,s»<0we have Euler's equation in spherical polar coordinates :
dv dv 1 dp
— + v— = .
dt dr p dr
(1)
The equation of continuity gives
rh) = F(t),
(2)
where F(t) is an arbitrary function of time; this equation expresses the fact that, since the
fluid is incompressible, the volume flowing through any spherical surface is independent of
the radius of that surface.
Fig. 3
Fig. 4
Substituting v from (2) in (1), we have
F'(t) dv 1 dp
— — + v — =
8r
p 8r
Integrating this equation over r from the instantaneous radius R = R(t) < a of the hole to
infinity, we obtain
F'(t) po
R P
(3)
where V = dR(t)!dt is the rate of change of the radius of the hole, and p is the pressure at
§10 Incompressible fluids 29
infinity; the fluid velocity at infinity is zero, and so is the pressure at the surface of the hole.
From equation (2) for points on the surface of the hole we find
F{i) = RHt)V{t\
and, substituting this expression for F(t) in (3), we obtain the equation
3F2 dF2 p
Integrating with the boundary condition V = for R — a (the fluid being initially at rest),
we have
dR _
Hence we have for the required total time for the hole to be filled
dt V [_ 3p \R3 /.
/ 3 P r dR
T ~ J 2^} ^/[{ajRf\]
This integral reduces to a beta function, and we have finally
V 2p Q r(i/3) V po
Problem 8. A sphere immersed in an incompressible fluid expands according to a given
law R = R(t). Determine the fluid pressure at the surface of the sphere.
Solution. Let the required pressure be P(t). Calculations exactly similar to those of
Problem 7, except that the pressure at r = R is P(t) and not zero, give instead of (3) the
equation
R * p p
and accordingly instead of (4) the equation
P 2 dR'
Bearing in mind the fact that V = dR/dt, we can write the expression for P(t) in the form
i rd2(i?2) /cLR\2
Problem 9. Determine the form of a jet emerging from an infinitely long slit in a plane
wall.
Solution. Let the wall be along the *axis in the xyplane, and the aperture be the
segment — \a < x < \a of that axis, the fluid occupying the halfplane y > 0. Far from the
wall [y *■ co) the fluid velocity is zero, and the pressure is p , say.
At the free surface of the jet (BC and B'C in Fig. 5a) the pressure p = 0, while the velocity
30
Ideal Fluids
§10
takes the constant value v x = V(2po/p)> by Bernoulli's equation. The wall lines are stream
lines, and continue into the free boundary of the jet. Let ip be zero on the line ABC; then,
on the line A'B'C, tft = —Q/p, where Q = pa^ is the rate at which the fluid emerges in
the jet (a lt v x being the jet width and velocity at infinity). The potential <f> varies from — oo
to + oo both along ABC and along A'B'C; let <f> be zero at B and B'. Then, in the plane of
the complex variable w, the region of flow is an infinite strip of width Qjp (Fig. 5b). (The
points in Fig. 5b, c, d are named to correspond with those in Fig. 5a.)
c,c
8
(c)
®
>o 3 £
I* ®
©
C
(b)
lQ/p
\B
<£ / B' Cjc' B A
=*1  "l r
(d)
Fig. 5
We introduce a new complex variable, the logarithm of the complex velocity:
f 1 da; "] v±
(1)
here v 1 e iin is the complex velocity of the jet at infinity. On A'B' we have 6 = 0; on AB,
9 ~ —t; on BC and B'C, v = v u while at infinity in the jet = \ir. In the plane of the
complex variable £, therefore, the region of flow is a semiinfinite strip of width n in the
right halfplane (Fig. 5c). If we can now find a conformal transformation which carries the
strip in the toplane into the halfstrip in the £plane (with the points corresponding as in
Fig. 5), we shall have determined w as a function of dzo/ds, and zv can then be found by a
simple quadrature.
In order to find the desired transformation, we introduce one further auxiliary complex
variable, u, such that the region of flow in the MpIane is the upper halfplane, the points
B and B' corresponding to u = ±1, the points C and C" to u — 0, and the infinitely distant
points A and A' to u = ± oo (Fig. 5d). The dependence of v) on this auxiliary variable is
given by the conformal transformation which carries the upper half of the uplane into the
strip in the wplane. With the above correspondence of points, this transformation is
to = lOgtt.
fm
(2)
In order to find the dependence of £ on u, we have to find a conformal transformation of the
halfstrip in the £plane into the upper half of the uplane. Regarding this halfstrip as a
§11 The drag force in potential flow past a body 31
triangle with one vertex at infinity, we can find the desired transformation by means of the
wellknown SchwarzChristoffel formula; it is
£ = — *sin 1 #. (3)
Formulae (2) and (3) give the solution of the problem, since they furnish the dependence of
dzo/dz on to in parametric form.
Let us now determine the form of the jet. On BC we have to — <f>, £ = i($n+0), while u
varies from 1 to 0. From (2) and (3) we obtain
<f> = ^log(cos0), (4)
prr
and from (1) we have
d(f}jdz = vie i0 ,
or
d* = dx + i dy = — e^ d<f> = V tanfl dd,
Vl IT
whence we find, by integration with the conditions y — 0,x = ia for 9 — —n, the form of the
jet, expressed parametrically. In particular, the compression of the jet is aja = 7r/(2+w)
= 061.
§11. The drag force in potential flow past a body
Let us consider the problem of potential flow of an incompressible ideal
fluid past some solid body. This problem is, of course, completely equivalent
to that of the motion of a fluid when the same body moves through it. To
obtain the latter case from the former, we need only change to a system of
coordinates in which the fluid is at rest at infinity. We shall, in fact, say in
what follows that the body is moving through the fluid.
Let us determine the nature of the fluid velocity distribution at great
distances from the moving body. The potential flow of an incompressible
fluid satisfies Laplace's equation, /\<f> = 0. We have to consider solutions
of this equation which vanish at infinity, since the fluid is at rest there.
We take the origin somewhere inside the moving body; the coordinate
system moves with the body, but we shall consider the fluid velocity distri
bution at a particular instant. As we know, Laplace's equation has a solution
1/r, where r is the distance from the origin. The gradient and higher space
derivatives of 1/r are also solutions. All these solutions, and any linear
combination of them, vanish at infinity. Hence the general form of the
required solution of Laplace's equation at great distances from the body is
a 1
<f> = h A«grad + ... ,
r r
where a and A are independent of the coordinates; the omitted terms con
tain higherorder derivatives of 1/r. It is easy to see that the constant a
must be zero. For the potential <f> = —ajr gives a velocity
v = — grad(a/r) = ar/r s .
Let us calculate the corresponding mass flux through some closed surface,
32 Ideal Fluids §11
say a sphere of radius R. On this surface the velocity is constant and equal
to a/R 2 ; the total flux through it is therefore p{a\R 2 )^R 2 = Am pa. But the
flux of an incompressible fluid through any closed surface must, of course,
be zero. Hence we conclude that a = 0.
Thus <j> contains terms of order \jr 2 and higher. Since we are seeking the
velocity at large distances, the terms of higher order may be neglected, and
we have
<f> = Agrad(l/r) = An/r 2 , (11.1)
and the velocity v = grad <f> is
v = (Agrad) grad = ^ ' , (11.2)
r r 3
where n is a unit vector in the direction of r. We see that at large distances
the velocity diminishes as l/r3. The vector A depends on the actual shape
and velocity of the body, and can be determined only by solving completely
the equation A<f> = at all distances, taking into account the appropriate
boundary conditions at the surface of the moving body.
The vector A which appears in (11.2) is related in a definite manner to
the total momentum and energy of the fluid in its motion past the body.
The total kinetic energy of the fluid (the internal energy of an incompressible
fluid is constant) is E = % fpv 2 dV, where the integration is taken over all
space outside the body. We take a region of space V bounded by a sphere of
large radius R, whose centre is at the origin, and first integrate only over
V, later letting R tend to infinity. We have identically
jv 2 dV = ju 2 dV+ j (v+u).(vu)dV,
where u is the velocity of the body. Since u is independent of the coordinates,
the first integral on the right is simply u 2 (V V ), where V is the volume of
the body. In the second integral, we write the sum v+uas grad (<f> + u . r) ;
using the facts that div v = (equation of continuity) and div u = 0, we
have
j* v 2 dV = u\V Vo)+ j div [(<£ + ur)(v u)]dF.
The second integral is now transformed into an integral over the surface S
of the sphere and the surface So of the body :
jv 2 dV = u 2 (VV )+ j> (^+u.r)(vu).df.
s+s a
On the surface of the body, the normal components of v and u are equal by
virtue of the boundary conditions ; since the vector df is along the normal
§11 The drag force in potential flow past a body 33
to the surface, it is clear that the integral over So vanishes identically. On
the remote surface S we substitute the expressions (11.1), (11.2) for <f> and v,
and neglect terms which vanish as R > oo. Writing the surface element
on the sphere S in the form df = nR 2 do, where do is an element of solid angle,
we obtain
jv*dV = u?(±ttB?Vo)+ J" [3(A.n)(u.n)(u.n)2fl3]do.
Finally, effecting the integration! and multiplying by \p> we obtain the
following expression for the total energy of the fluid :
E = £ p (4ttA.u V u*). (11.3)
As has been mentioned already, the exact calculation of the vector A
requires a complete solution of the equation /\<f> = 0» taking into account the
particular boundary conditions at the surface of the body. However, the
general nature of the dependence of A on the velocity u of the body can be
found directly from the facts that the equation is linear in <f>, and the boundary
conditions are linear in both <f> and u. It follows from this that A must be a
linear function of the components of u. The energy E given by formula
(11.3) is therefore a quadratic function of the components of u, and can be
written in the form
E = \m ik UiU k , (11.4)
where m^ is some constant symmetrical tensor, whose components can be
calculated from those of A; it is called the inducedmass tensor.
Knowing the energy E, we can obtain an expression for the total momentum
P of the fluid. To do so, we notice that infinitesimal changes in E and P
are related by J dE = u • dP; it follows from this that, if E is expressed in
t The integration over o is equivalent to averaging the integrand over all directions of the vector
n and multiplying by 4w. To average expressions of t he ty pe (A • n)(B • n) = ^4,« tJ BfcWfc, where A, B
are constant vectors, we notice that the mean values tiittje form a symmetrical tensor, which can be
expressed in terms of the unit tensor S«: w,w* = aS,*. Contracting with respect to the suffixes i
and k, and remembering that «*»!< = 1, we find that a = \. Hence
(AnXB.n) = VfaAtB* = £A B.
X For, let the body be accelerated by some external force F. The momentum of the fluid will thereby
be increased; let it increase by dP during a time df. This increase is related to the force by dP = F df,
and on scalar multiplication by the velocity u we have u • dP = F • u df, i.e. the work done by the
force F acting through the distance u df , which in turn must be equal to the increase dE in the energy
of the fluid.
It should be noticed that it would not be possible to calculate the momentum directly as the integral
pv dV over the whole volume of the fluid. The reason is that this integral, with the velocity v
distributed in accordance with (11.2), diverges, in the sense that the result of the integration, though
finite, depends on how the integral is taken: on effecting the integration over a large region, whose
dimensions subsequently tend to infinity, we obtain a value depending on the shape of the region
(sphere, cylinder, etc.). The method of calculating the momentum which we use here, starting from
the relation u • dP = dE, leads to a completely definite final result, given by formula (11.6), which
certainly satisfies the physical relation between the rate of change of the momentum and the forces
acting on the body.
34 Ideal Fluids §n
the form (11.4), the components of P must be
Pi = m ik u k . (H.5)
Finally, a comparison of formulae (11.3), (11.4) and (11.5) shows that P
is given in terms of A by
P = 4tt P A p Vo\i. (11.6)
It must be noticed that the total momentum of the fluid is a perfectly definite
finite quantity.
The momentum transmitted to the fluid by the body in unit time is dP/d*.
With the opposite sign it evidently gives the reaction F of the fluid, i.e. the
force acting on the body :
F = dP/d*. (H.7)
The component of F parallel to the velocity of the body is called the drag
force, and the perpendicular component is called the lift force.
If it were possible to have potential flow past a body moving uniformly
in an ideal fluid, we should have P = constant, since u = constant, and so
F = 0. That is, there would be no drag and no lift; the pressure forces
exerted on the body by the fluid would balance out (a result known as
d'Alemberfs paradox). The origin of this paradox is most clearly seen by
considering the drag. The presence of a drag force in uniform motion of a
body would mean that, to maintain the motion, work must be continually
done by some external force, this work being either dissipated in the fluid or
converted into kinetic energy of the fluid, and the result being a continual
flow of energy to infinity in the fluid. There is, however, by definition
no dissipation of energy in an ideal fluid, and the velocity of the fluid set in
motion by the body diminishes so rapidly with increasing distance from the
body that there can be no flow of energy to infinity.
However, it must be emphasised that all these arguments relate only to
the motion of a body in an infinite volume of fluid. If, for example, the
fluid has a free surface, a body moving uniformly parallel to this surface will
experience a drag. The appearance of this force (called wave drag) is due to
the occurrence of a system of waves propagated on the free surface, which
continually remove energy to infinity.
Suppose that a body is executing an oscillatory motion under the action
of an external force f. When the conditions discussed in §10 are fulfilled,
the fluid surrounding the body moves in a potential flow, and we can use the
relations previously obtained to derive the equations of motion of the body.
The force f must be equal to the time derivative of the total momentum of
the system, and the total momentum is the sum of the momentum Mvl
of the body (M being the mass of the body) and the momentum P of the fluid :
Mdu/d*+dP/d* = f.
§1 1 The drag force in potential flow past a body 35
Using (11.5), we then obtain
M dui/dt + mac dujc/dt = /*,
which can also be written
^(M8 ik + m ik )=f i . (11.8)
at
This is the equation of motion of a body immersed in an ideal fluid.
Let us now consider what is in some ways the converse problem. Suppose
that the fluid executes some oscillatory motion on account of some cause
external to the body. This motion will set the body in motion also.f We
shall derive the equation of motion of the body.
We assume that the velocity of the fluid varies only slightly over distances
of the order of the dimension of the body. Let v be what the fluid velocity
at the position of the body would be if the body were absent; that is, v is the
velocity of the unperturbed flow. According to the above assumption, v
may be supposed constant throughout the volume occupied by the body.
We denote the velocity of the body by u as before.
The force which acts on the body and sets it in motion can be determined
as follows. If the body were wholly carried along with the fluid (i.e. if
v = u), the force acting on it would be the same as the force which would act
on the liquid in the same volume if the body were absent. The momentum of
this volume of fluid is pVo\, and therefore the force on it is pVo dv/d*.
In reality, however, the body is not wholly carried along with the fluid;
there is a motion of the body relative to the fluid, in consequence of which
the fluid itself acquires some additional motion. The resulting additional
momentum of the fluid is mi k {u k v k ), since in (11.5) we must now replace u
by the velocity uv of the body relative to the fluid. The change in this
momentum with time results in the appearance of an additional reaction
force on the body of m ik d{u k v k )\dt. Thus the total force on the body is
pVo— m ik —(u k v k ).
dt dt
This force is to be equated to the time derivative of the body momentum.
Thus we obtain the following equation of motion:
d dvi d
—(Mui) = pV — m ik —(u k v k ).
dt dt dt
Integrating both sides with respect to time, we have
Mui = pVoVim ik (u k v k ),
or
(MS ik + m ik )u k = (m ik + pVo8i k )v k . (11.9)
f For example, we may be considering the motion of a body in a fluid through which a sound wave
is propagated, the wavelength being large compared with the dimension of the body.
36 Ideal Fluids
§12
We put the constant of integration equal to zero, since the velocity u of
the body in its motion caused by the fluid must vanish when v vanishes.
The relation obtained determines the velocity of the body from that of the
fluid. If the density of the body is equal to that of the fluid (M = P V ),
we have u = v, as we should expect.
PROBLEMS
sin°p UT i? N * ? omparin S f (" 1 ) ^th the expression for * for flow past a sphere obtained in
$1U, Problem 2, we see that
where R is the radius of the sphere. The total momentum transmitted to the fluid bv the
sphere is, according to (11.6), P = firpJPu, so that the tensor m t1c is
• P™.™ 1 . Obtain i the equation of motion for a sphere executing an oscillatory motion
in an ideal fluid, and for a sphere set in motion by an oscillating fluid.
he expression i
A = IRSvl,
The total mon
pF?u, so that
Wik = frrpR 3 8nc.
The drag on the moving sphere is
F = frrpl&du/dt,
and the equation of motion of the sphere oscillating in the fluid is
i^ipo+ip)^. = f,
where Po is the density of the sphere. The coefficient of du/dtis the virtual mass of the sphere •
it consists of the actual mass of the sphere and the induced mass, which in this case is half
the mass of the fluid displaced by the sphere.
If the sphere is set in motion by the fluid, we have for its velocity, from (1 1.9),
3p
u = v.
p + 2po
If the density of the sphere exceeds that of the fluid ( Po >/>),«< v, i.e. the sphere "lags
behind the fluid; if Po < p, on the other hand, the sphere "goes ahead".
Problem 2. Express the moment of the forces acting on a body moving in a fluid in
terms of the vector A.
Solution As we know from mechanics, the moment M of the forces acting on a body is
iT™£ sa T lt8 e /f«ran«Wtt function (in this case, the energy E) by the relation
7 . ' d ' where S0 ls the vector of an infinitesimal rotation of the body, and &E is the
resulting change in E. Instead of rotating the body through an angle W (and correspondingly
changing the components mic), we may rotate the fluid through an angle 80 relative to the
body (and correspondingly change the velocity u). We have Su = — 80 xu, so that
8E= PSu = SSuxP.
Using the expression (11.6) for P, we then obtain the required formula:
M = uxP = 4tt P A x u.
§12. Gravity waves
The free surface of a liquid in equilibrium in a gravitational field is a plane.
If, under the action of some external perturbation, the surface is moved
§12 Gravity waves 37
from its equilibrium position at some point, motion will occur in the liquid.
This motion will be propagated over the whole surface in the form of waves,
which are called gravity waves, since they are due to the action of the gravita
tional field. Gravity waves appear mainly on the surface of the liquid,
they affect the interior also, but less and less at greater and greater depths.
We shall here consider gravity waves in which the velocity of the moving
fluid particles is so small that we may neglect the term (v • grad)v in compari
son with dvjdt in Euler's equation. The physical significance of this is easily
seen. During a time interval of the order of the period t of the oscillations
of the fluid particles in the wave, these particles travel a distance of the order
of the amplitude a of the wave. Their velocity is therefore of the order of
afr. It varies noticeably over time intervals of the order of t and distances
of the order of A in the direction of propagation (where A is the wavelength).
Hence the time derivative of the velocity is of the order of v/t, and the space
derivatives are of the order of vj\. Thus the condition (v • grad)v <^ dvjdt
is equivalent to
1 (a\ 2 a \
A \t/ t r
or
a < A, (12.1)
i.e. the amplitude of the oscillations in the wave must be small compared with
the wavelength. We have seen in §9 that, if the term (v«grad)v in the
equation of motion may be neglected, we have potential flow. Assuming the
fluid incompressible, we can therefore use equations (10.6) and (10.7).
The term \v 2 in the latter equation may be neglected, since it contains the
square of the velocity; putting f(t) = and including a term pgz on account
of the gravitational field, we obtain
P = pgzpd+Jdt. (12.2)
We take the sraxis vertically upwards, as usual, and the ryplane in the
equilibrium surface of the fluid.
Let us denote by £ the z coordinate of a point on the surface; £ is a func
tion of x, y and t. In equilibrium £ = 0, so that £ gives the vertical displace
ment of the surface in its oscillations. Let a constant pressure po (for example,
the atmospheric pressure) act on the surface. Then we have at the surface,
by (12.2),
Po = —pgl — ptyjdt.
Instead of the potential <j>, we can use a potential <j>' = </> + (pojp)t\ this makes
no difference, since v = grad <f> = grad <f>'. The term p is removed from
the above equation, however, and on dropping the prime we obtain the
condition at the surface as
gt+(d<f>jdt) z ^ = 0. (12.3)
Since the amplitude of the wave oscillations is small, the displacement I
38 Ideal Fluids §12
is small. Hence we can suppose, to the same degree of approximation, that
the vertical component of the velocity of points on the surface is simply the
time derivative of £:
But v z = 8<f>/8z, so that
W/9*)zt = dt/dt.
Substituting £ from (12.3) we have
\8z gdfl/^
Since the oscillations are small, we can take the value of the parenthesis
at z = instead of z = £. Thus we have finally the following system of
equations to determine the motion in a gravitational field:
A<f> = 0, (12.4)
(86 1 8U\
h^dr =° (12.5)
\8z g 8t 2 7 Z=0 v }
We shall here consider waves on the surface of a fluid whose area is un
limited, and we shall also suppose that the wavelength is small in comparison
with the depth of the fluid; we can then regard the fluid as infinitely deep.
We shall therefore omit the boundary conditions at the sides and bottom.
Let us consider a gravity wave propagated along the #axis and uniform
in the ^direction; in such a wave, all quantities are independent of y. We
shall seek a solution which is a simple periodic function of time and of the
coordinate x, i.e. we put
<f> = f(z) cos (kx— cot).
Here <o is what is called the circular frequency (we shall say simply the
frequency) of the wave; lnjoa is the period of the motion at a given point;
k is called the wave number; A = 2njk is the wavelength, i.e. the period of the
motion along the xaxis at a given time.
Substituting in the equation
8U 8%
A<^ = — + — = 0,
8x 2 8z 2
we have
d 2 //d#2#2f = o.
This equation has the solutions e ks and e~ kz . We must take the former,
since the latter gives an unlimited increase of <f> as we go into the interior of
§12 Gravity waves 39
the fluid (we recall that the fluid occupies the region z < 0). Thus we obtain
for the velocity potential
<f> = Ae kz cos (kx cot). (12.6)
We have also to satisfy the boundary condition (12.5). Substituting (12.6),
we obtain
kafi/g = 0,
or
ft>2 = kg. (12.7)
This gives the relation between the wave number and the frequency of a
gravity wave.
The velocity distribution in the moving fluid is found by simply taking
the space derivatives of <f> :
v x =  Ake kz sin (kx  cot), v z = Ake kz cos (kx  cot). (12.8)
We see that the velocity diminishes exponentially as we go into the fluid.
At any given point in space (i.e. for given x, z) the velocity vector rotates
uniformly in the ##plane, its magnitude remaining constant and equal to
Ake hz .
Let us also determine the paths of fluid particles in the wave. We tem
porarily denote by x, z the coordinates of a moving fluid particle (and not
of a point fixed in space), and by *o, %o the values of re and z at the equilibrium
position of the particle. Then v x = d#/d*, v z = dz/dt, and on the right
hand side of (12.8) we may approximate by writing xo, zq in place of x, z,
since the oscillations are small. An integration with respect to time then
gives
k
x—xo= — A — e kz o cos (kxo — cot),
CO
k (12  9)
zzq =  A — e kz o sin (kx  cot).
CO
Thus the fluid particles describe circles of radius (Akjco)e h ^ about the points
(*o, #o); this radius diminishes exponentially with increasing depth.
The velocity of propagation U of the wave is, as we shall show in §66,
U = Bwjdk. Substituting here <o = \/(kg), we find that the velocity of pro
pagation of gravity waves on an unbounded surface of infinitely deep fluid
is
u = W(g/k) = M£A/27r). v i2.i0)
It increases with wavelength.
40 Ideal Fluids §12
PROBLEMS
Problem 1. Determine the velocity of propagation of gravity waves on an unbounded
surface of fluid of depth h.
Solution. At the bottom of the fluid, the normal velocity component must be zero,
i.e. v z = 8<f>ldz = f or z = —h. From this condition we find the ratio of the constants
A and B in the general solution
<f> = [Ae kz +Bete]cos(kxcot).
The result is
<f> = A cos(kx — cot) cosh. k(z+h).
From the boundary condition (12.5) we find the relation between k and to to be
co 2 — gk tanh kh.
The velocity of propagation of the wave is
1 / s r , ,, kh \
— — tanhM + .
V k tanh kh _ cosh 2 kh J
For M> 1 we have the result (12.10), and for kh < 1 the result (13.10) (see below).
u =
2
Problem 2. Determine the relation between frequency and wavelength for gravity waves
on the surface separating two fluids, the upper fluid being bounded above by a fixed horizontal
plane, and the lower fluid being similarly bounded below. The density and depth of the
lower fluid are p and h, those of the upper fluid are p' and h', and p > />'.
Solution. We take the «yplane as the equilibrium plane of separation of the two fluids.
Let us seek a solution having in the two fluids the forms
<f> = A cosh k{z+h) cos(kx— cot),
4>' = B cosh. k(z—h')cos(kx— cot),
so that the conditions at the upper and lower boundaries are satisfied; see the solution to
Problem 1. At the surface of separation, the pressure must be continuous; by (12.2), this
gives the condition
8<f> 8cf)'
Pgt+P— = Pgt+p'— for z = £,
ot ot
i = ^( p 'i^ t ) (2)
Moreover, the velocity component v z must be the same for each fluid at the surface of separa
tion. This gives the condition
8<f>/8z = dtfjdz for z = 0. (3)
Now Vz = 8<f>/8z ~ dtjet and, substituting (2), we have
86 8 2 <f>' 8 2 cf>
5(PP') = P'^^. (4)
Substituting (1) in (3) and (4) gives two homogeneous linear equations for A and B, and the
§12 Gravity waves 41
condition of compatibility gives
2 kg(pp')
CO* =
p coth.kh\p coihhh'
For kh^> 1, kh'^> 1 (both fluids very deep),
pp
,2 =
p + p
while for kh <^ 1, kh' <^ 1 (long waves),
g{pp')hh'
= k / g{ppy
V ph' + p',
ph' + p'h
Problem 3. Determine the relation between frequency and wavelength for gravity waves
propagated simultaneously on the surface of separation and on the upper surface of two fluid
layers, the lower (of density p) being infinitely deep, and the upper (of density p') being of
depth h' and having a free upper surface.
Solution. We take the aryplane as the equilibrium plane of separation of the two fluids.
Let us seek a solution having in the two fluids the forms
= Ae kz cos(&v cot),
(f)' = [Be~ kz +Ce kz ] cos(kx cot).
At the surface of separation, i.e. for z = 0, we have the conditions (see Problem 2)
8$ af 8* ,s*f a^
 = ^. S{PP}^ = P^^ (2)
and at the upper surface, i.e. for z = h', the condition
dz g dt z
The first equation (2), on substitution in (1), gives A — C—B, and the remaining two con
ditions then give two equations for B and C; from the condition of compatibility we obtain a
quadratic equation for o> 2 , whose roots are
CO" 2 = &£ , CO* = kg.
S p+ p ' + (pp')e2M> *
For h' *■ oo these roots correspond to waves propagated independently on the surface of
separation and on the upper surface.
Problem 4. Determine the possible frequencies of oscillationf (stationary waves) of a
fluid of depth h in a rectangular tank of width a and length b.
Solution. We take the * and y axes along two sides of the tank. Let us seek a solution
in the form of a stationary wave:
(f> = f(x,y) cosh k(z + h) cos cot.
t See §68.
42 Ideal Fluids §13
We obtain for /the equation
3 2 f 8 2 f
dx 2 dy 2
and the condition at the free surface gives, as in Problem 1, the relation
co 2 = gk tanh kh.
We take the solution of the equation for/ in the form
/ = cos/w cos qy, p 2 + q 2 = k 2 .
At the sides of the tank we must have the conditions
v x = d<t>jdx = for x = 0, a;
v y = 8<f>/dy = for y = 0, b.
Hence we find p — mirfa, q = nn/b, where m, n are integers. The possible values of k %
are therefore
k 2 = 7T 2 l — + —
U 2 b 2 J
§13. Long gravity waves
Having considered gravity waves of length small compared with the depth
of the fluid, let us now discuss the opposite limiting case of waves of length
large compared with the depth. These are called long waves.
Let us examine first the propagation of long waves in a channel. The
channel is supposed to be along the #axis, and of infinite length. The
crosssection of the channel may have any shape, and may vary along its
length. We denote the crosssectional area of the fluid in the channel by
S = S(x, t). The depth and width of the channel are supposed small in
comparison with the wavelength.
We shall here consider longitudinal waves, in which the fluid moves along
the channel. In such waves the velocity component v x along the channel is
large compared with the components v y , v z .
We denote v x by v simply, and omit small terms. The ^component of
Euler's equation can then be written in the form
d<v 1 dp
8t p dx
and the ^component in the form
1 dp
p dz
we omit terms quadratic in the velocity, since the amplitude of the wave is
§13 Long gravity waves 43
again supposed small. From the second equation we have, since the pressure
at the free surface (z = £) must be po,
P =#)+#>(£ 4
Substituting this expression in the first equation, we obtain
dvjdt = gdlfdx. (13.1)
The second equation needed to determine the two unknowns v and £ can
be derived similarly to the equation of continuity; it is essentially the equation
of continuity for the case in question. Let us consider a volume of fluid
bounded by two plane crosssections of the channel at a distance dx apart.
In unit time a volume (Sv) x of fluid flows through one plane, and a volume
(Sv) x +ax through the other. Hence the volume of fluid between the two
planes changes by
8(Sv)
(Sv) x+Ax (Sv) x = — — dx.
ox
Since the fluid is incompressible, however, this change must be due simply to
the change in the level of the fluid. The change per unit time in the volume
of fluid between the two planes considered is (dSJdt)dx. We can therefore
write
8S 8(Sv)
dx = dx,
dt dx
or
dS 8(Sv)
— + —  = 0. (13.2)
8t dx
This is the required equation of continuity.
Let .So be the equilibrium crosssectional area of the fluid in the channel.
Then S = So+S', where S' is the change in the crosssectional area caused
by the wave. Since the change in the fluid level is small, we can write S'
in the form ££, where b is the width of the channel at the surface of the fluid.
Equation (13.2) then becomes
81 8(S v)
b± + JlAjL = o. (13.3)
8t 8x
Differentiating (13.3) with respect to t and substituting 8v\8t from (13.1),
we obtain
8H g d I 8l\ _
~8&~bl)x\ °!hc) ~ ^ ' '
If the channel crosssection is the same at all points, then So = constant
and
8H gS e 2 £
1 _ *_!? _± = o. (13.5)
8& b 8x 2 v J
This is called a wave equation: as we shall show in §63, it corresponds to
44 Ideal Fluids §14
the propagation of waves with a velocity U which is independent of frequency
and is the square root of the coefficient of 8 2 £/8x 2 . Thus the velocity of propa
gation of long gravity waves in channels is
U = V(gSo/b). (13.6)
In an entirely similar manner, we can consider long waves in a large tank,
which we suppose infinite in two directions (those of x and y). The depth
of fluid in the tank is denoted by h. The component v z of the velocity is now
small. Euler's equations take a form similar to (13.1):
8v x dl dv y 81
The equation of continuity is derived in the same way as (13.2) and is
8h 8{hv x ) dihvy)
8t dx 8y
We write the depth h as h +£, where h is the equilibrium depth. Then
$t d(hov x ) 8(hov y )
Let us assume that the tank has a horizontal bottom (ho = constant).
Differentiating (13.8) with respect to t and substituting (13.7), we obtain
TF gh [l* + w) ' (13 ' 9)
This is again a (twodimensional) wave equation; it corresponds to waves
propagated with a velocity
U = V(gfy (13.10)
§14. Waves in an incompressible fluid
There is a kind of gravity wave which can be propagated inside an incom
pressible fluid. Such waves are due to an inhomogeneity of the fluid caused
by the gravitational field. The pressure (and therefore the entropy s) neces
sarily varies with height; hence any displacement of a fluid particle in height
destroys the mechanical equilibrium, and consequently causes an oscillatory
motion. For, since the motion is adiabatic, the particle carries with it to its
new position its old entropy s, which is not the same as the equilibrium value
at the new position.
We shall suppose below that the wavelength is small in comparison with
distances over which the gravitational field causes a marked change in density;
and we shall regard the fluid itself as incompressible. This means that we
can neglect the change in its density caused by the pressure change in the
§14 Waves in an incompressible fluid 45
wave. The change in density caused by thermal expansion cannot be neglec
ted, since it is this that causes the phenomenon in question.
Let us write down a system of hydrodynamic equations for this motion.
We shall use a suffix to distinguish the values of quantities in mechanical
equilibrium, and a prime to mark small deviations from those values. Then
the equation of conservation of the entropy s = sq+s' can be written, to
the first order of smallness,
ds'ldt+vgradso = 0, (14.1)
where so, like the equilibrium values of other quantities, is a given function
of the vertical coordinate z.
Next, in Euler's equation we again neglect the term (v • grad)v (since
the oscillations are small); taking into account also the fact that the equili
brium pressure distribution is given by grad po = pog, we have to the same
accuracy
dv grad/) grad/>' grad/>o ,
St p po p 2
Since, from what has been said above, the change in density is due only to
the change in entropy, and not to the change in pressure, we can put
\ &sq fp
and we then obtain Euler's equation in the form
*.«(*•) ,_**£. (14.2)
dt po\ dso fp po
We can take po under the gradient operator, since, as stated above, we always
neglect the change in the equilibrium density over distances of the order of a
wavelength. The density may likewise be supposed constant in the equation
of continuity, which then becomes
div v = 0. (14.3)
We shall seek a solution of equations (141)— (14.3) in the form of a plane
wave:
v = constant x e Uk  r ~ ut \
and similarly for s' and />'. Substitution in the equation of continuity (14.3)
gives
vk = 0, (14.4)
i.e. the fluid velocity is everywhere perpendicular to the wave vector k (a
transverse wave). Equations (14.1) and (14.2) give
teas —
1 / 8 P0 \ , ik ,
= v«gradso> —iosw = — I I s g p .
po \ «fro 1 v po
46 Ideal Fluids §14
The condition v • k = gives with the second of these equations
** (£)/■*
and, eliminating v and s' from the two equations, we obtain the desired
relation between the wave vector and the frequency,
1 / dp \ ds
Here and henceforward we omit the suffix zero to the equilibrium values of
thermodynamic quantities; the #axis is vertically upwards, and 6 is the
angle between this axis and the direction of k. If the expression on the right
of (14.5) is positive, the condition for the stability of the equilibrium distribu
tion s(z) (the condition that convection is absent — see §4) is fulfilled.
We see that the frequency depends only on the direction of the wave
vector, and not on its magnitude. For 6 = we have w = 0; this means
that waves of the type considered, with the wave vector vertical, cannot
exist.
If the fluid is in both mechanical equilibrium and complete thermodynamic
equilibrium, its temperature is constant and we can write
^1 _ / 8s \ d P _ ( 6s \
dz~ \dp) T dz Pg \dpJ T '
Finally, using the wellknown thermodynamic relations
\8p) T p\dT) p * \ds) P CpKdTjp
where c p is the specific heat per unit mass, we find
/ T g I dp \
In particular, for a perfect gas,
m ~vbo siaS  < i4  7 >
CHAPTER II
VISCOUS FLUIDS
§15. The equations of motion of a viscous fluid
Let us now study the effect of energy dissipation, occurring during the
motion of a fluid, on that motion itself. This process is the result of the
thermodynamic irreversibility of the motion. This irreversibility always
occurs to some extent, and is due to internal friction (viscosity) and thermal
conduction.
In order to obtain the equations describing the motion of a viscous fluid,
we have to include some additional terms in the equation of motion of an ideal
fluid. The equation of continuity, as we see from its derivation, is equally
valid for any fluid, whether viscous or not. Euler's equation, on the other
hand, requires modification.
We have seen in §7 that Euler's equation can be written in the form
dt dxic
where II <* is the momentum flux density tensor. The momentum flux
given by formula (7.2) represents a completely reversible transfer of momen
tum, due simply to the mechanical transport of the different particles of fluid
from place to place and to the pressure forces acting in the fluid. The viscosity
(internal friction) is due to another, irreversible, transfer of momentum from
points where the velocity is large to those where it is small.
The equation of motion of a viscous fluid may therefore be obtained by
adding to the "ideal" momentum flux (7.2) a term — o'ac which gives the
irreversible "viscous" transfer of momentum in the fluid. Thus we write
the momentum flux density tensor in a viscous fluid in the form
Ilifc = pSik+pviVjca'ik = — Pffc+pOiZfe. (15.1)
The tensor
<*ik = —pbik+o'ik (15.2)
is called the stress tensor, and o'tk the viscosity stress tensor, 0% gives the part
of the momentum flux that is not due to the direct transfer of momentum
with the mass of moving fluid.f
The general form of the tensor a'ik can be established as follows. Processes
f We shall see below that a' a contains a term proportional to Sa, i.e. of the same form as the
term pSa. When the momentum flux tensor is put in such a form, therefore, we should specify what
is meant by the pressure p; see the end of §49.
47
48 Viscous Fluids §15
of internal friction occur in a fluid only when different fluid particles move
with different velocities, so that there is a relative motion between various
parts of the fluid. Hence a'ijc must depend on the space derivatives of the
velocity. If the velocity gradients are small, we may suppose that the momen
tum transfer due to viscosity depends only on the first derivatives of the velo
city. To the same approximation, o'oc may be supposed a linear function of
the derivatives dvt/dxk. There can be no terms in o'ac independent of
dvijdxjc, since a'ik must vanish for v — constant. Next, we notice that o'tk
must also vanish when the whole fluid is in uniform rotation, since it is clear
that in such a motion no internal friction occurs in the fluid. In uniform rota
tion with angular velocity SI, the velocity v is equal to the vector product
ftxr. The sums
8vt dvjc
dxjc Sxt
are linear combinations of the derivatives dvi/dxic, and vanish when v = Slxr.
Hence o'ik must contain just these symmetrical combinations of the deriva
tives dvi/dxjc.
The most general tensor of rank two satisfying the above conditions is
/ dvt 8v k \ 8v t
a a = a — + —  +b—8 ik ,
\ OXjc OXi / oxi
where a and b are independent of the velocity.f It is convenient, however,
to write this expression in a slightly different form, in which a and b are
replaced by other constants:
o i* = t] — + —  8«r— 1 +£8o—. (15.3)
\ OXjc OXi OX\ I OXl
The expression in parentheses has the property of vanishing on contraction
with respect to i and k. The constants r\ and £ are called coefficients of viscosity.
As we shall show in §§16 and 49, they are both positive:
•n > 0, £ > 0. (15.4)
The equations of motion of a viscous fluid can now be obtained by simply
adding the expressions da'ujdxie to the righthand side of Euler's equation
/ 8vt dvi \ dp
P\ 1" V]c = —
*\ 8t dxjcf
8xi
Thus we have
/ dvi dvi \
P h Vic
\ 8t dxjcf
dp 8 1/ 8vi 8vjc 8vi\\ 8 / dv t \
=  — + — U — + — P**— ) + — U— • (15.5)
8xi 8xjc \ \ dxjc 8xi 8xi J ) dxt \ 8xi J
f In malang this statement we use the fact that the fluid is isotropic, as a result of which its proper
ties must be described by scalar quantities only (in this case, a and b).
§15 The equations of motion of a viscous fluid 49
This is the most general form of the equations of motion of a viscous fluid.
The quantities rj and £ are functions of pressure and temperature. In
general, p and T, and therefore rj and £, are not constant throughout the
fluid, so that v\ and £ cannot be taken outside the gradient operator.
In most cases, however, the viscosity coefficients do not change noticeably
in the fluid, and they may be regarded as constant. We then have
Sff'a; _ / &Vi 8 8v k 2 d dvi\ a dvi
dxjc \ dxjcdxic dxi dxjc 3 dx\ dxi / dx% dxi
8 2 Vi d dvi
OXlJOXic OXi OXi
But
dvi/dxi = divv, d^i/dx^xjc == A»<
Hence we can write the equation of motion of a viscous fluid, in vector form,
p\ f (v«grad)v = grad/> + i7Av + (£ + ^)graddivv. C15.6)
If the fluid may be regarded as incompressible, div v = 0, and the last
term on the right of (15.6) is zero. Thus the equation of motion of an
incompressible viscous fluid is
dv 1 7)
h (v«grad)v = gradp + A v. (15.7)
dt p p
This is called the NavierStokes equation. The stress tensor in an incom
pressible fluid takes the simple form
(dv{ dvjc \
h— — I. (15.8)
CXlc OXi I
We see that the viscosity of an incompressible fluid is determined by only
one coefficient. Since most fluids may be regarded as practically incompres
sible, it is this viscosity coefficient t] which is generally of importance. The
ratio
v = vIp (15.9)
is called the kinematic viscosity (while rj itself is called the dynamic viscosity).
We give below the values of ?) and v for various fluids, at a temperature of
20° C:
r\ (g/cm sec) v (cm 2 /sec)
Water 0010 0010
Air 000018 0150
Alcohol 0018 0022
Glycerine 85 68
Mercury 00156 00012
50 Viscous Fluids §15
It may be mentioned that the dynamic viscosity of a gas at a given tempera
ture is independent of the pressure. The kinematic viscosity, however, is
inversely proportional to the pressure.
The pressure can be eliminated from the NavierStokes equation in the
same way as from Euler's equation. Taking the curl of both sides of equation
(15.7), we obtain, instead of equation (2.11) as for an ideal fluid,
8
— (curl v) = curl (vx curl v) + vA(curlv). (15.10)
We must also write down the boundary conditions on the equations of
motion of a viscous fluid. There are always forces of molecular attraction
between a viscous fluid and the surface of a solid body, and these forces have
the result that the layer of fluid immediately adjacent to the surface is brought
completely to rest, and "adheres" to the surface. Accordingly, the boundary
conditions on the equations of motion of a viscous fluid require that the fluid
velocity should vanish at fixed solid surfaces:
v = 0. (15.11)
It should be emphasised that both the normal and the tangential velocity
component must vanish, whereas for an ideal fluid the boundary conditions
require only the vanishing of v n .f
In the general case of a moving surface, the velocity v must be equal to
the velocity of the surface.
It is easy to write down an expression for the force acting on a solid
surface bounding the fluid. The force acting on an element of the surface is
just the momentum flux through this element. The momentum flux through
the surface element df is
nW/& = (pvivjc— oi]c)6f}c.
Writing &f k in the form d/& = tik df, where n is a unit vector along the normal,
and recalling that v = at a solid surface, $ we find that the force P acting
on unit surface area is
Pi = OiWUc = pnia' ik n k . (15.12)
The first term is the ordinary pressure of the fluid, while the second is the
force of friction, due to the viscosity, acting on the surface. We must em
phasise that n in (15.12) is a unit vector along the outward normal to the fluid,
i.e. along the inward normal to the solid surface.
If we have a surface of separation between two immiscible fluids, the
conditions at the surface are that the velocities of the fluids must be equal
f We may note that, in general, Euler's equations cannot be satisfied with the boundary condi
tion v = 0.
X In determining the force acting on the surface, each surface element must be considered in a
frame of reference in which it is at rest. The force is equal to the momentum flux only when the
surface is fixed.
§15 The equations of motion of a viscous fluid 51
and the forces which they exert on each other must be equal and opposite.
The latter condition is written
«l,fc oi,tt + «2,t 02, tk = 0,
where the suffixes 1 and 2 refer to the two fluids. The normal vectors ni and
112 are in opposite directions, i.e. »i f < = — »2,* = »*, so that we can write
»< 01, ik = ni ai,ik (15.13)
At a free surface of the fluid the condition
(Jikfik = a'iknic—pni = (15.14)
must hold.
We give below, for reference, expressions for the components of the stress
tensor and the NavierStokes equation in cylindrical and spherical co
ordinates. In cylindrical coordinates r, <f>, z the components of the stress
tensor are
8v r /l 8v r 8v^ V 6 \
^=^ + 2,—, ^ = ^__ + ___j,
l\ dvt v r \ I 8v$ 1 8v z \
8v z I 8v z 8v r \
a. ,+2^, ,„.„(_ + _). (15 .15)
The three components of the NavierStokes equation and the equation of
continuity are
8v r 8v r v 6 8v r 8v r v s 2
+ v r + — —  + v z —
8t 8r r 8(f> 8z r
\8p t 8 2 v r 1 8 2 v r 8 2 v r 1 8v r 2 8v$ v r \
p 8r \ 8r 2 ~r* 8<f> 2 8z 2 r~8~r ~~ ~^~8~I~ "r 2 )'
8V4 8v^ v* dv, to* v^
8t 8r r 8(f> 8z r
1 dp /8 2 v 6 1 8 2 v# 8*vt 1 ty 2 dvr v+\
pr 8<f> \8r 2 r 2 8<f> 2 8z 2 r 8r r 2 8<f> r 2 V
8v z 8v z Vx 8v z 8v z
1 v r h — —  + v z
8t 8r r 8$ 8z
1 8p / 8 2 v z 1 8 2 v z 8 2 v z 1 8v z \
= \ v \ 1 1 1 I
P 8z \ 8r 2 r 2 8<j> 2 8z 2 r 8r V
8v r 1 dv+ 8v z v r ,,*
— +  + — + _ = o. (15.16)
8r r 8j> 8z r J
(15.17)
52 Viscous Fluids §15
In spherical coordinates r, <£, 9 we have for the stress tensor
8v r
8r
°M = P + 2 V ( — — — + — + ,
\ r sin 8<p r r ]
~ V \r 86 + ~8r~~7) y
"" ^Usin^ 8<f> r 89 r F
(dVj 1 d*, r w .\
\ dr r sm 9 8<f> r I
while the equations of motion are
8v r 8v r v 8v r v A 8v r v^+vJ
1 v r 1 1  L.
8t dr r 86 rsin9 8<f> r
\8p rl
= + v \
p 8r lr
Oaa = — ■
<JrO
\8p pi 8 2 (rv r ) 1 8 2 v r 1 8 2 v r cot 6 8v r
p 8r lr 8r 2 r 2 86* r 2 sin 2 9 8<f> 2 r 2 89
2 8v e
8vo_ 2 8v 4 2v r 2cot0 1
89 ~ r 2 sin9^~~r^ r^~ T
8v 8v e v d dv e v* 8v v r v e v A 2 cot6
+ v r 1 1 ^ j. " *
8t 8r r 89 rsin0 8(f> r
— d JL J 1 8 ^ rv ^ 1 g % 1 8 2 v cot9 8v e
pr 89 V [r 8r 2 r 2 89* r 2 sin 2 9 8<j> 2 + r 2 ~8~9
2 cos 9 8v$ 2 8v r v e
r 2 sin 2 9 8<f> r 2 89 r 2 sin 2 9
]•
8v 4> , dv <i> , *>0 &># V, 8v, VrV, V e V,COt9
8t 8r r 89 r sin9 8<f> r r
^__J g/> ri ^ 2 K) i a% l a%
pr sin 9 8<f> V [r 8r 2 r 2 80 2 r 2 sin 2 9 8<j? +
cot 9 8v 4> 2 8v r 2cos0 8v e v^
r 2 89 r 2 sin9 8<f> r 2 sin 2 9 8<f> r 2 sin 2 9
]■
8v r 1 8v e 1 8v & 2v r v g cot9
— +  ~ + —^~ a ~ + — + = 0. (15.18)
8r r 89 rsm9 8j> r r J
§16 Energy dissipation in an incompressible fluid 53
Finally, we give the equation that must be satisfied by the stream function
*fj(x,y) in twodimensional flow of an incompressible viscous fluid. It is
obtained by substituting v x = difj/dy, v y = 8tf,]dx, v z = in equation
(15.10):
8 (ai\ # g(A»A) , H g( A 0)
^ Alfl) Y x ^ + Vy^r v ^ = ()  (15 ' 19)
§16. Energy dissipation in an incompressible fluid
The presence of viscosity results in the dissipation of energy, which is
finally transformed into heat. The calculation of the energy dissipation is
especially simple for an incompressible fluid.
The total kinetic energy of an incompressible fluid is
£kin = pjVdF.
We take the time derivative of this energy, writing 8(%pv 2 )ldt = pvtdvijdt
and substituting for 8v\\8t the expression for it given by the NavierStokes
equation :
dvi _ dvi 1 dp 1 da'ik
The result is
8t dxjc p 8x% p dxjc
d 1 da'
— (w v2 ) = pv«(vgrad)vv«grad/>+^
dt ° dx k
=  / ,(v.grad)(i^ + ^) + div(v.a')cT^— .
\ p I oxk
Here v • a' denotes the vector whose components are Vio'uc Since div v =
for an incompressible fluid, we can write the first term on the right as a
divergence :
(1^2) = _ div ^ v ^ 2 + t^j _ v . <J o'i^. (16.1)
The expression in brackets is just the energy flux density in the fluid:
the term pv$v 2 +pjp) is the energy flux due to the actual transfer of fluid
mass, and is the same as the energy flux in an ideal fluid (see (10.5)). The
second term, v • o\ is the energy flux due to processes of internal friction.
For the presence of viscosity results in a momentum flux a' a; a transfer
of momentum, however, always involves a transfer of energy, and the
energy flux is clearly equal to the scalar product of the momentum flux
and the velocity.
54 Viscous Fluids §16
If we integrate (16.1) over some volume V, we obtain
d
Yt
U P v*dV=  L Lv/i^ + ^v.o'j.df ( a' ik — dV. (16.2)
The first term on the right gives the rate of change of the kinetic energy of
the fluid in V owing to the energy flux through the surface bounding V.
The integral in the second term is consequently the decrease per unit time
in the kinetic energy owing to dissipation.
If the integration is extended to the whole volume of the fluid, the surface
integral vanishes (since the velocity vanishes at infinity]), and we find the
energy dissipated per unit time in the whole fluid to be
■Ekin = — cr'oc dV.
J dxjc
In incompressible fluids, the tensor cr'a is given by (15.8), so that
dvi dvi 1 dvi dvjc \
a ik = 7] ( 1 J.
dxjc 8x k \ dxjc 8xi 1
It is easy to verify that this expression can be written
/ 8vj dv k \ 2
\ dx k dxt I
Thus we have finally for the energy dissipation in an incompressible fluid
• , C ( dvi dv k \ 2
^HU + id dF  (16  3)
The < dissipation leads to a decrease in the mechanical energy, i.e. we must
have E kia < 0. The integral in (16.3), however, is always positive. We
therefore conclude that the viscosity coefficient rj is always positive.
PROBLEM
Transform the integral (16.3) for potential flow into an integral over the surface bounding
the region of flow.
Solution. Putting dv t /8xk — dvkldxt and integrating once by parts, we find
^= 2 "J(£) v = 2 "J£^
■E'kin = — t) J grada 2 «df.
t We are considering the motion of the fluid in a system of coordinates such that the fluid is at
rest at infinity. Here, and in similar cases, we speak, for the sake of definiteness, of an infinite volume
of fluid, but this implies no loss of generality. For a fluid enclosed in a finite volume, the surface
integral again vanishes, because the normal velocity component at the surface vanishes.
§17 FUm in a pipe 55
§17. Flow in a pipe
We shall now consider some simple problems of motion of an incom
pressible viscous fluid.
Let the fluid be enclosed between two parallel planes moving with a
constant relative velocity u. We take one of these planes as the #;srplane,
with the #axis in the direction of u. It is clear that all quantities depend only
on y, and that the fluid velocity is everywhere in the ^direction. We have
from (15.7) for steady flow
dpjdy = 0, d 2 v/dy 2 = 0.
(The equation of continuity is satisfied identically.) Hence p = constant,
v = ay+b. For y — and y = h{h being the distance between the planes)
we must have respectively v = and v = u. Thus
v = yu/h. (17.1)
The fluid velocity distribution is therefore linear. The mean fluid velocity,
defined as
1 h
V = h J V dy '
is
v = \u. (17.2)
From (15.12) we find that the normal component of the force on either plane
is just p, as it should be, while the tangential friction force on the plane
y = is
v xy = q dv/dy = rju/h; (17.3)
the force on the plane y = h is — rjujk.
Next, let us consider steady flow between two fixed parallel planes in the
presence of a pressure gradient. We choose the coordinates as before;
the ffaxis is in the direction of motion of the fluid. The NavierStokes
equations give, since the velocity clearly depends only on y,
d 2 v 1 dp dp
dy 2 r] 8x dy
The second equation shows that the pressure is independent of y, i.e. it
is constant across the depth of the fluid between the planes. The righthand
side of the first equation is therefore a function of x only, while the lefthand
side is a function of y only; this can be true only if both sides are constant.
Thus dpjdx — constant, i.e. the pressure is a linear function of the coordi
nate x along the direction of flow. For the velocity we now obtain
1 dp
v = — — y 2 + ay+b.
2?7 dx
56 Viscous Fluids §17
The constants a and b are determined from the boundary conditions, v =
for y = and jy = h. The result is
°=^Si A2 0vW]. (17.4)
Thus the velocity varies parabolically across the fluid, reaching its maximum
value in the middle. The mean fluid velocity (averaged over the depth of the
fluid) is again
1 h
on calculating this, we find
h* dp
We may also calculate the frictional force a xy = r}(dv/dy)y= acting on one
of the fixed planes. Substitution from (17.4) gives
o xy = \h dp/dx. (17.6)
Finally, let us consider steady flow in a pipe of arbitrary crosssection
(the same along the whole length of the pipe, however). We take the axis of
the pipe as the ffaxis. The fluid velocity is evidently along the *axis at all
points, and is a function of y and z only. The equation of continuity is
satisfied identically, while the y and z components of the NavierStokes
equation again give dpjdy = dp/dz = 0, i.e. the pressure is constant over
the crosssection of the pipe. The ^component of equation (15.7) gives
d"h) 8 z v 1 dp
Qy2 Q z 2 yj dx
Hence we again conclude that dp/dx = constant; the pressure gradient may
therefore be written  Ap//, where Ap is the pressure difference between the
ends of the pipe and / is its length.
Thus the velocity distribution for flow in a pipe is determined by a two
dimensional equation of the form A v = constant. This equation has to be
solved with the boundary condition v = at the circumference of the cross
section of the pipe. We shall solve the equation for a pipe of circular cross
section. Taking the origin at the centre of the circle and using polar co
ordinates, we have by symmetry v = v(r). Using the expression for the
Laplacian in polar coordinates, we have
1 d / dv\ Ap
r dr \ dr J rjl
§17 Flow in a pipe 57
Integrating, we find
Ap
v= r 2 + alogr + b. (17.8)
The constant a must be put equal to zero, since the velocity must remain
finite at the centre of the pipe. The constant b is determined from the re
quirement that v = for r = R, where R is the radius of the pipe. We then
find
v = L{R2r*). (17.9)
Thus the velocity distribution across the pipe is parabolic.
It is easy to determine the mass Q of fluid passing each second through
any crosssection of the pipe (called the discharge). A mass p • 2tttv dV
passes each second through an annular element litr dr of the crosssectional
area. Hence
R
Q = 2np I rvdr.
Using (17.9), we obtain
ttA/>
Q = £T&. (17.10)
The mass of fluid is thus proportional to the fourth power of the radius of the
pipe (Poiseuille's formula).
PROBLEMS
Problem 1. Determine the flow in a pipe of annular crosssection, the internal and external
radii being R lt i? 8 .
Solution. Determining the constants a and b in the general solution (17.8) from the con
ditions that v = for r — R t and r = R tt we find
V =
A M„, • . R*Ri*
r R 2 *Ri* r 1
lR 2 2_ r 2 + — t —log— .
L log(i? 2 /i?i) g £ 2 J
The discharge is
■nAp [ (R 2 2 Ri 2 ) 2 !
Q = — — #2 4 #i 4  — —
^ 8v L loRiRtlRi) J
log(Rt/Ri)
Problem 2. The same as Problem 1, but for a pipe of elliptical crosssection.
Solution. We seek a solution of equation (17.7) in the form v = Ay i +Bz a +C. The
constants A, B, C axe determined from the requirement that this expression must satisfy
the boundary condition » = 0on the circumference of the ellipse (i.e. Ay i +Bz i +C =
58 Viscous Fluids §17
must be the same as the equation y*la*+z*/b 2 = 1, where a and b are the semiaxes of the
ellipse). The result is
Ap a 2 b 2
V =
The discharge is
Q =
2r;l a*
irAp a%^
4vl a* + b*
Problem 3. The same as Problem 1, but for a pipe whose crosssection is an equilateral
triangle of side a.
Solution. The solution of equation (17.7) which vanishes on the bounding triangle is
Ap 2
where h u ha, h a are the lengths of the perpendiculars from a given point in the triangle to its
three sides. For each of the expressions A/»i, AAj, M» (where A = 8 2 l8x a + 8 2 l8y 2 ) is
zero; this is seen at once from the fact that each of the perpendiculars h lt h t , h a may be taken
as the axis of y or z, and the result of applying the Laplacian to a coordinate is zero. We
therefore have
A(fah2h 3 ) = 2{hi grad V grad A 3 +/t 2 grad A 3 . grad Ai +
+fa grad Ar grad A 2 ).
But gradAj  n», gradA 2 = n 2 , gradAj = n S) where n^ n 2 , n s are unit vectors along
the perpendiculars h lt h 2t h a . Any two of n lt n s , n 8 are at an angle 2tt/3, so that grad h x . grad hi
= ninj = cos (2tt/3) = — £, and so on. We thus obtain the relation
Aihfahs) = {hi + h 2 + hz) = JV3«,
and we see that equation (17.7) is satisfied. The discharge is
V3a 4 A/>
=
32(M
Problem 4. A cylinder of radius i? x moves with velocity u inside a coaxial cylinder of
radius R it their axes being parallel. Determine the motion of a fluid occupying the space
between the cylinders.
Solution. We take cylindrical coordinates, with the zaxis along the axis of the cylinders.
The velocity is everywhere along the sraxis and depends only on r (as does the pressure):
v z — v(r). We obtain for v the equation
1 d / dv\
r dr\ dr/
the term (vgrad)v = v 8v/8z vanishes identically. Using the boundary conditions v = u
for r — JRj and v = for r — R 2 , we find
log(r/U a )
v = u
logCRi/lfc)
The frictional force per unit length of either cylinder is 277r^//log(i? 2 / J R 1 ).
§17 Flow in a pipe 59
Problem 5. A layer of fluid of thickness h is bounded above by a free surface and below
by a fixed plane inclined at an angle a to the horizontal. Determine the flow due to gravity.
Solution. We take the fixed plane as the *yplane, with the *axis in the direction of
flow (Fig. 6). We seek a solution depending only on z. The NavierStokes equations with
Vx — v{z) in a gravitational field are
d 2 v dp
r)— + pg sin* = 0, — + p£cosa = 0.
dz 2 dz
At the free surface (z = h) we must have a xt = ydv/dz = 0, a zt = — p = —po (.Po being
the atmospheric pressure). For z = we must have v = 0. The solution satisfying these
conditions is
pg sin a
p = po + pg(h  z) cos a, v = — z(2h  z).
It]
The discharge, per unit length in the ydirection, is
pgh 3 sin a
vdz = .
= pJ
Fig. 6
Problem 6. Determine the way in which the pressure falls along a tube of circular cross
section in which a viscous perfect gas is flowing isothermally (bearing in mind that the
dynamic viscosity t\ of a perfect gas is independent of the pressure).
Solution. Over any short section of the pipe the gas may be supposed incompressible,
provided that the pressure gradient is not too great, and we can therefore use formula
(17.10), according to which
<fr = SrjQ
dx TTpR*
Over greater distances, however, p varies, and the pressure is not a linear function of *.
According to the equation of state, the gas density p = mplkT, where m is the mass of a
molecule and k is Boltzmann's constant, so that
dp _ SrjQkT 1
dx irmR* p
(The discharge Q of the gas through the tube is obviously the same, whether or not the gas
is incompressible.) From this we find
where p2, pi are the pressures at the ends of a section of the tube of length /.
60 Viscous Fluids §18
§18. Flow between rotating cylinders
Let us now consider the motion of a fluid between two infinite coaxial
cylinders of radii R lf R 2 (R 2 > i?i), rotating about their axis with angular
velocities Q lf Q2. We take cylindrical coordinates r, <f>, z, with the saxis
along the axis of the cylinders. It is evident from symmetry that
v z = v r = 0, V4 = v(r), p = p(r).
The NavierStokes equation in cylindrical coordinates gives in this case two
equations :
dp/dr = pv 2 /r, (18.1)
d 2 v 1 dv v
dr 2 r dr r 2
The latter equation has solutions of the form r n \ substitution gives n = ± 1,
so that
b
v = ar + .
r
The constants a and b are found from the boundary conditions, according to
which the fluid velocity at the inner and outer cylindrical surfaces must be
equal to that of the corresponding cylinder: v = Ri&i for r = i?i, v = R2O.2
for r = R 2 . As a result we find the velocity distribution to be
n 2 ^2 2 ^i^ 1 2 (niQ 2 )#iW 1
" " RfRf ' + WRP ? (18 ' 3 >
The pressure distribution is then found from (18.1) by straightforward
integration.
For Qi = &2 = & we have simply v = Dr, i.e. the fluid rotates rigidly
with the cylinders. When the outer cylinder is absent (D.2 = 0, R 2 = 00)
we have v = D.iRi 2 /r.
Let us also determine the moment of the frictional forces acting on the
cylinders. The frictional force acting on unit area of the inner cylinder is
along the tangent to the surface and, from (15.12), is equal to the component
a' r $ of the stress tensor. Using formulae (15.15), we find
o (Q 1 Q 2 )R 2 2
— —2rt .
' RJRf
The force acting on unit length of the cylinder is obtained by multiplying
§19 The law of similarity 61
by 2irRi, and the moment M\ of that force by multiplying the result by R\.
We thus have
WQiQ 2 )fliW
* *■*. ■ (18 ' 4)
The moment M2 of the forces acting on the inner cylinder is clearly — Mi.f
The following general remark may be made concerning the solutions of the
equations of motion of a viscous fluid which we have obtained in §§17 and 18.
In all these cases the nonlinear term (v • grad)v in the equations which
determine the velocity distribution is identically zero, so that we are actually
solving linear equations, a fact which very much simplifies the problem.
For this reason all the solutions also satisfy the equations of motion for an
incompressible ideal fluid, say in the form (10.2) and (10.3). This is why
formulae (17.1) and (18.3) do not contain the viscosity coefficient at all.
This coefficient appears only in formulae, such as (17.9), which relate the
velocity to the pressure gradient in the fluid, since the presence of a pressure
gradient is due to the viscosity; an ideal fluid could flow in a pipe even if
there were no pressure gradient.
§19. The law of similarity
In studying the motion of viscous fluids we can obtain a number of impor
tant results from simple arguments concerning the dimensions of various
physical quantities. Let us consider any particular type of motion, for
instance the motion of a body of some definite shape through a fluid. If the
body is not a sphere, its direction of motion must also be specified : e.g. the
motion of an ellipsoid in the direction of its greatest or least axis. Alternatively,
we may be considering flow in a region with boundaries of a definite form
(a pipe of given crosssection, etc.).
In such a case we say that bodies of the same shape are geometrically similar;
they can be obtained from one another by changing all linear dimensions in
the same ratio. Hence, if the shape of the body is given, it suffices to specify
any one of its linear dimensions (the radius of a sphere or of a cylindrical
pipe, one semiaxis of a spheroid of given eccentricity, and so on) in order
to determine its dimensions completely.
We shall at present consider steady flow. If, for example, we are discussing
flow past a solid body (which case we shall take below, for definiteness), the
velocity of the main stream must therefore be constant. We shall suppose
the fluid incompressible.
Of the parameters which characterise the fluid itself, only the kinematic
f The solution of the more complex problem of the motion of a viscous fluid in a narrow space
between cylinders whose axes are parallel but not coincident may be found in: N. E. Kochin, I. A.
Kibel' and N. V. Roze, Theoretical Hydromechanics (Teoreticheskaya gidromekhanika), Part 2, 3rd
ed., p. 419, Moscow 1948; A. Sommerfeld, Mechanics of Deformable Bodies, §36, Academic Press,
New York 1950.
62 ' Viscous Fluids §19
viscosity v = rj/p appears in the equations of hydrodynamics (the Navier
Stokes equations); the unknown functions which have to be determined by
solving the equations are the velocity v and the ratio pip of the pressure
p to the constant density p. Moreover, the flow depends, through the
boundary conditions, on the shape and dimensions of the body moving
through the fluid and on its velocity. Since the shape of the body is supposed
given, its geometrical properties are determined by one linear dimension,
which we denote by /. Let the velocity of the main stream be u. Then any
flow is specified by three parameters, v, u and /. These quantities have the
following dimensions :
v = cm 2 /sec, / = cm, u = cm/sec.
It is easy to verify that only one dimensionless quantity can be formed from
the above three, namely uljv. This combination is called the Reynolds
number and is denoted by R:
R = puljrj = uljv. (19.1)
Any other dimensionless parameter can be written as a function of R.
We shall now measure lengths in terms of /, and velocities in terms of u,
i.e. we introduce the dimensionless quantities r//, v/w. Since the only
dimensionless parameter is the Reynolds number, it is evident that the velocity
distribution obtained by solving the equations of incompressible flow is
given by a function of the form
v = «f(r//, R). (19.2)
It is seen from this expression that, in two different flows of the same type
(for example, flow past spheres of different radii by fluids of different vis
cosities), the velocities vju are the same functions of the ratio x\l if the Reynolds
number is the same for each flow. Flows which can be obtained from
one another by simply changing the unit of measurement of coordinates and
velocities are said to be similar. Thus flows of the same type with the same
Reynolds number are similar. This is called the law of similarity (O. Rey
nolds 1883).
A formula similar to (19.2) can be written for the pressure distribution in
the fluid. To do so, we must construct from the parameters v y I, u some
quantity with the dimensions of pressure divided by density; this quantity
can be w 2 , for example. Then we can say that p/pu 2 is a function of the dimen
sionless variable r// and the dimensionless parameter R. Thus
p = /o« 2 /(r//, R). (19.3)
Finally, similar considerations can also be applied to quantities which
characterise the flow but are not functions of the coordinates. Such a
quantity is, for instance, the drag force F acting on the body. We can say
that the dimensionless ratio of F to some quantity formed from v, «, /, p
§20 Stokes' formula 63
and having the dimensions of force must be a function of the Reynolds num
ber alone. Such a combination of v> u, I, p can be pu 2 l 2 , for example. Then
F = pu*Pf(R). (19.4)
If the force of gravity has an important effect on the flow, then the latter
is determined not by three but by four parameters, /, u, v and the acceleration
g due to gravity. From these parameters we can construct not one but two
independent dimensionless quantities. These can be, for instance, the Rey
nolds number and the Froude number, which is
F = u*llg. (19.5)
In formulae (19.2)(19.4) the function /will now depend on not one but two
parameters (R and F), and two flows will be similar only if both these num
bers have the same values.
Finally, we may say a little regarding nonsteady flows. A nonsteady flow
of a given type is characterised not only by the quantities v, w, / but also
by some time interval t characteristic of the flow, which determines the rate
of change of the flow. For instance, in oscillations, according to a given law,
of a solid body, of a given shape, immersed in a fluid, t may be the period of
oscillation. From the four quantities v, u, /, r we can again construct two
independent dimensionless quantities, which may be the Reynolds number
and the number
S = ut/1, (19.6)
sometimes called the Strouhal number. Similar motion takes place in these
cases only if both these numbers have the same values.
If the oscillations of the fluid occur spontaneously (and not under the action
of a given external exciting force), then for motion of a given type S will be
a definite function of R:
S=/(R).
§20. Stokes' formula
The NavierStokes equation is considerably simplified in the case of flow
at small Reynolds numbers. For steady flow of an incompressible fluid, this
equation is
(v.grad)v= (l//>)grad/> + (V/°)Av.
The term (v • grad)v is of the order of magnitude of w 2 //, u and / having the
same meaning as in §19. The quantity (rj/p) Av is of the order of magnitude
of rju[pl 2 . The ratio of the two is just the Reynolds number. Hence the term
(v • grad)v may be neglected if the Reynolds number is small, and the
equation of motion reduces to a linear equation
*?Avgrad/> = 0. (20.1)
64 Viscous Fluids §20
Together with the equation of continuity
div v = (20.2)
it completely determines the motion. It is useful to note also the equation
A curlv = 0, (20.3)
which is obtained by taking the curl of equation (20.1).
As an example, let us consider rectilinear and uniform motion of a sphere
in a viscous fluid. The problem of the motion of a sphere, it is clear, is
exactly equivalent to that of flow past a fixed sphere, the fluid having a
given velocity u at infinity. The velocity distribution in the first problem is
obtained from that in the second problem by simply subtracting the velocity
u; the fluid is then at rest at infinity, while the sphere moves with velocity
u. If we regard the flow as steady, we must, of course, speak of the flow
past a fixed sphere, since, when the sphere moves, the velocity of the fluid at
any point in space varies with time.
Thus we must have v = u at infinity; we write v = v'f u, so that v'
is zero at infinity. Since div v = div v' = 0, v' can be written as the curl of
some vector : v = curl A+ u. The curl of a polar vector is well known to be
an axial vector, and vice versa. Since the velocity is an ordinary polar vector,
A must be an axial vector. Now v, and therefore A, depend only on the radius
vector r (we take the origin at the centre of the sphere) and on the parameter
u; both these vectors are polar. Furthermore, A must evidently be a linear
function of u. The only such axial vector which can be constructed for a
completely symmetrical body (the sphere) from two polar vectors is the
vector product rxu. Hence A must be of the form /'(r)nxu, where f(r)
is a scalar function of r, and n is a unit vector in the direction of the radius
vector. The product f(r)n can be written as the gradient, grad /(»), of some
function /(r), so that the general form of A is grad/xu. Hence we can write
the velocity v' as
v' = curl [grad/xu].
Since u is a constant, grad/xu = curl(/u), so that
v = curl curl (/u) + u. (20.4)
To determine the function /, we use equation (20.3). Since
curlv = curl curl curl(/u) = (grad div A)curl(/u)
= A curl(/u),
(20.3) takes the form A 2 curl (Ju) = 0, or, since u = constant,
A 2 (grad/xu) = (A 2 grad/)xu = 0.
It follows from this that
A 2 grad/ = 0. (20.5)
§20 Stokes' formula 65
A first integration gives
A 2 / = constant.
It is easy to see that the constant must be zero, since the velocity v must
vanish at infinity, and so must its derivatives. The expression A 2 / contains
fourth derivatives of /, whilst the velocity is given in terms of the second
derivatives of/. Thus we have
1 d
A 2 /=
Hence
1 d / d \
r* dr\ or J
A/= 2a/r + A.
The constant A must be zero if the velocity is to vanish at infinity. From
A/ = 2a/r we obtain
f=ar+b/r. (20.6)
The additive constant is omitted, since it is immaterial (the velocity being
given by derivatives of/).
Substituting in (20.4), we have after a simple calculation
u+n(u«n) 3n(u«n)u
v = Ma   + b . (20.7)
r r 3
The constants a and b have to be determined from the boundary conditions :
at the surface of the sphere (r = R), v = 0, i.e.
u + n(u»n) 3n(u«n) — u
\xa + b — = 0.
R R?
Since this equation must hold for all n, the coefficients of u and n(u • n)
must each vanish :
a b a 3b
— + — 1 = 0, + — = 0.
R R? R R?
Hence a = f i?, b = %R 3 . Thus we have finally
f=lRr+lR?lr, (20.8)
„ u + n(u«n) u3n(u«n)
v =  f # i L  IR* S L + u> (20.9)
r r 3
or, in spherical components,
3R R3
r 3R ft*!
V r = U COS 1 1 ,
L 2r 2r3j'
. A r 3R R3]
g — —u sin 1 .
9 L 4r 4r3j
(20.10)
66 Viscous Fluids
§20
This gives the velocity distribution about the moving sphere. To determine
the pressure, we substitute (20.4) in (20.1):
gradp = qAv = yj A curl curl (/u)
= V A (grad div (/u)  u A/).
But A 2 / =0, and so
gradp = grad[7jAdiv(/u)] = gradfrugrad A/).
Hence
/> = rju . grad Af+po, (20. 1 1)
where />o is the fluid pressure at infinity. Substitution for /leads to the final
expression
u*n
P = Po  h^R. (20.12)
Using the above formulae, we can calculate the force F exerted on the
sphere by the moving fluid (or, what is the same thing, the drag on the sphere
as it moves through the fluid). To do so, we take spherical coordinates with
the polar axis parallel to u; by symmetry, all quantities are functions only of
r and of the polar angle 0. The force F is evidently parallel to the velocity u.
The magnitude of this force can be determined from (15.12). Taking from
this formula the components, normal and tangential to the surface, of the
force on an element of the surface of the sphere, and projecting these compo
nents on the direction of u, we find
F = j>(p cos 6+ a'rr cos  a rd sin 0)d/, (20. 13)
where the integration is taken over the whole surface of the sphere.
Substituting the expressions (20.10) in the formulae
dr \r 86 dr r /
(see (15.17)), we find that at the surface of the sphere
o'rr = 0, a' re =  (3r)/2R)u sin 0,
while the pressure (20.12) is p = p (3 V l2R)u cos 0. Hence the integral
(20.13) reduces to F = (3?/w/2#) § d/, or, finally,f
F = 67TRr)u. (20.14)
This formula (called Stokes' formula) gives the drag on a sphere moving
f f T^on ^7 to u som e later applications, we may mention that, if the calculations are done with
formula (20.7) for the velocity (the constants a and b being undetermined), we find
F = 87rarju. (20.14a)
§20 Stokes' formula 67
slowly in a fluid. We may notice that the drag is proportional to the first
powers of the velocity and linear dimension of the body.f
This dependence of the drag on the velocity and dimension holds for
slowlymoving bodies of other shapes also. The direction of the drag on a
body of arbitrary shape is not the same as that of the velocity; the general
form of the dependence of F on u can be written
Ft = flan* (20.15)
where aue is a tensor of rank two, independent of the velocity. It is important
to note that this tensor is symmetrical {am — aid), a result which holds in the
linear approximation with respect to the velocity, and is a particular case of
a general law valid for slow motion accompanied by dissipative processes.!
The solution that we have just obtained for flow past a sphere is not
valid at great distances from it, even if the Reynolds number is small. In
order to see this, we estimate the magnitude of the term (v • grad)v, which
we neglected in (20.1). At great distances the velocity is u. The derivatives
of the velocity at these distances are seen from (20.9) to be of the order of
uR/r 2 . Thus (v • grad)v is of the order of u 2 R/r z . The terms retained in
equation (20.1), for example (l//>) grad/>, are of the order r]Rujpr z (cf. (20.12)).
The condition
uqR/prZ > uWlr 2
holds only at distances r <^ vju, where v = rjjp. At greater distances, the
terms we have omitted cannot legitimately be neglected, and the velocity
distribution obtained is incorrect.
To obtain the velocity distribution at great distances from the body,
we have to take into account the term (vgrad)v omitted in (20.1). Since
the velocity v is nearly equal to u at these distances, we can put approximately
U'grad in place of v»grad. We then find for the velocity at great distances
the linear equation
(ugrad)v = (l//o) grad/> + vAv (20.16)
(C. W. Oseen, 1910).
We shall not pause to give here the solution of this equation for flow
f The drag can also be calculated for a slowlymoving ellipsoid of any shape. The corresponding
formulae are given by H. Lamb, Hydrodynamics, 6th ed., §339, Cambridge 1932. We give here the
limiting expressions for a plane circular disk of radius R moving perpendicular to its plane :
F = \(yrjRu
and for a similar disk moving in its plane:
F = 32r)Ruj3.
X See, for instance, Statistical Physics, §120, Pergamon Press, London 1958.
6% Viscous Fluids
§20
past a sphere,f but merely mention that the velocity distribution thus
obtained can be used to derive a more accurate formula for the drag on the
sphere, which includes the next term in the expansion of the drag in powers
of the Reynolds number uRJv. This formula is J
F = (m^uRl 1 + _ I. (20.17)
Finally, we may mention that, in solving the problem of flow past an
infinite cylinder with the main stream perpendicular to the axis of the
cylinder, Oseen's equation has to be used from the start; in this case, equation
(20.1) has no solution which satisfies the boundary conditions at the surface
of the cylinder and at the same time vanishes at infinity. The drag per unit
length of the cylinder is found to be
4tttjU
= hylog(uR/4 v y < 20  18 )
where y = 0577 is Euler's constant.
PROBLEMS
Problem 1 Determine the motion of a fluid occupying the space between two concentric
spheres of radii R lt R z (R 2 > RJ, rotating uniformly about different diameters with angular
velocities R lt fi 2 ; the Reynolds numbers <W/", iW/" are small compared with unity.
Solution On account of the linearity of the equations, the motion between two rotating
spheres may be regarded as a superposition of the two motions obtained when one sphere is
at rest and the other rotates. We first put fi a = 0, i.e. only the inner sphere is rotating. It
is reasonable to suppose that the fluid velocity at every point is along the tangent to a circle
m a plane perpendicular to the axis of rotation with its centre on the axis. On account of
the axial symmetry, the pressure gradient in this direction is zero. Hence the equation of
motion (20.1) becomes Av = 0. The angular velocity vector H, is an axial vector Argu
ments similar to those given previously show that the velocity can be written as
v = curl[/(r)fti] = grad/x fli.
The equation of motion then gives grad A/X Si x = 0. Since the vector grad A / is parallel
to the radius vector, and the vector product rXfi, cannot be zero for given Si t and arbitrarv
r, we must have grad A/ = 0, so that
A/= constant.
\ A I / ietai , Ied account of the calculations for a sphere and a cylinder is given by N E Kochin
I. A. Kibel and N. V. Rozb, Theoretical Hydromechanics (Teoreticheskaya gidromekhanika), Part 2,
brid C e 1932 aPtei " §§25 ~ 26, Moscow 1948 5 H  Lamb, Hydrodynamics, 6th ed., §§3423, Cam
X At first sight it might appear that Osben's equation, which does not correctly give the velocity
distribution near the sphere, could not be used to calculate the correction to the drag. In fact however
the contribution to F due to the motion of the neighbouring fluid (where u < vlr) must be expanded in
powers of the vector u. The first nonzero correction term in F arising from this contribution is
then proportional to « 2 u, i.e. is of the second order with respect to the Reynolds number: it therefore
does not affect the firstorder correction in formula (20.17). Further corrections to Stokes' formula
cannot be calculated from Oseen's formula.
§20 Stokes' formula 69
Integrating, we find
f=arZ + , v= ( — 2afiixr.
The constants a and 6 are found from the conditions that v = for r = R 2 and v = u
for r = R u where u = Si x xr is the velocity of points on the rotating sphere. The result is
_ fli3J?2 3 / 1 1 \
#2 3 #l 3 I 7» " 5^ j ^ X r "
The fluid pressure is constant (p = p ). Similarly, we have for the case where the outer
sphere rotates and the inner one is at rest (i^ = 0)
V =
R 2 sR
\P = Po)
i is at res
R^R 2 3
8/1 1 \
» 1 3\ j R 1 3 r 3)
R 2 3 Ri
In the general case where both spheres rotate, we have
V =
RM 2 3
'*[(* h)* xt+ (hh)* Xi
R 2 *R
If the outer sphere is absent (R 2 = 00, Q 2 = 0), i.e. we have simply a sphere of radius R
rotating in an infinite fluid, then
V = (#3/ r 3) Sl xr
Let us calculate the moment of the frictional forces acting on the sphere in this case. If we
take spherical coordinates with the polar axis parallel to SI, we have v r = v & = 0, v 6 — v
= (R 3 £l/r 2 ) sin 0. The frictional force on unit area of the sphere is
/ / 8v v\
a r 4> = 7][ =— 3r]Q, sin 6.
The total moment on the sphere is
M = j o' H Rsmd 2ttR? sin 66,
whence we find
M= SirrjRSQ.
If the inner sphere is absent, v = £l 2 X r, i.e. the fluid simply rotates rigidly with the sphere
surrounding it.
Problem 2. Determine the velocity of a spherical drop of fluid (of viscosity •>?') moving
under gravity in a fluid of viscosity 17 (W. Rybczynski 1911).
Solution. We use a system of coordinates in which the drop is at rest. For the fluid
outside the drop we again seek a solution of equation (20.5) in the form (20.6), so that the
velocity has the form (20.7). For the fluid inside the drop, we have to find a solution which
does not have a singularity at r = (and the second derivatives of/, which determine the
velocity, must also remain finite). This solution is
70 Viscous Fluids §20
and the corresponding velocity is
v = ^u4£r 2 [n(u.n)2u].
At the surface of the sphere t the following conditions must be satisfied. The normal velocity
components outside (vg) and inside (v<) the drop must be zero:
Vi,r = Ve t r = 0.
The tangential velocity component must be continuous:
Vl,d = Ve,6>
as must be the component a T $ of the stress tensor :
&i,r8 = a e,ro
The condition that the stress tensor components a rr are equal need not be written down;
it would determine the required velocity u, which is more simply found in the manner shown
below. From the above four conditions we obtain four equations for the constants a, b,A, B,
whose solutions are
a = R 2ri + 3ri ' . t = R > 1' .. . AB* V
4(l+l')' 4{,+V)' 2(1+1')
By (20.14a), we have for the drag
F = 2TTWt]R{h] + 3r)')/(ri + r)').
As i\ *■ oo Ccorresponding to a solid sphere) this formula becomes Stokes' formula. In the
limit q' *■ (corresponding to a gas bubble) we have F = AtrwqR, i.e. the drag is twothirds
of that on a solid sphere.
Equating F to the force of gravity on the drop, %nR z {p — p')g, we find
2R ^{pp')(r ) + rj')
u = .
3rj(2rj + 3r)')
Problem 3. Two parallel plane circular disks (of radius R) lie one above the other a small
distance apart; the space between them is filled with fluid. The disks approach at a constant
velocity «, displacing the fluid. Determine the resistance to their motion (O. Reynolds).
Solution. We take cylindrical coordinates, with the origin at the centre of the lower disk,
which we suppose fixed. The flow is axisymmetric and, since the fluid layer is thin, pre
dominantly radial: v z <^.v r , and also 8v r /8r <^.8v T /dz. Hence the equations of motion
become
d*v r dp 8p
1 dz* 8r dz '
lM + ^ = 0) (2)
r dr dz
f We may neglect the change of shape of the drop in its motion, since this change is of a higher
order of smallness. However, it must be borne in mind that, in order that the moving drop should
in fact be spherical, the forces due to surface tension at its boundary must exceed the forces due to
pressure differences, which tend to make the drop nonspherical. This means that we must have
•qujR <^ a.jR, where a is the surfacetension coefficient, or, substituting u ~ R^gpjq,
R < V(*!pg)
§21 The laminar wake 71
with the boundary conditions
at* =
at z = h
atr = R
v r — v z = 0;
v r = 0, v z = —u\
P = po,
where h is the distance between the disks, and po the external pressure. From equations (1)
we find
1 dp
Integrating equation (2) with respect to z, we obtain
1 d r A3 a / dp \
u = I rv r az = \r — I,
rdrj \Znrdr\&)
o
whence
3inu
P=Po + ^(R 2 r*).
The total resistance to the moving disk is
F = 3irr)uR*/2hK
§21. The laminar wake
In steady flow of a viscous fluid past a solid body, the flow at great distances
behind the body has certain characteristics which can be investigated inde
pendently of the particular shape of the body.
Let us denote by U the constant velocity of the incident current; we take
the direction of U as the *axis, with the origin somewhere inside the body.
The actual fluid velocity at any point may be written U+v; v vanishes at
infinity.
It is found that, at great distances behind the body, the velocity v is
noticeably different from zero only in a relatively narrow region near the
#axis. This region, called the laminar zvake,f is reached by fluid particles
which move along streamlines passing fairly close to the body. Hence the
flow in the wake is essentially rotational. On the other hand, the viscosity has
almost no effect at any point on streamlines that do not pass near the body,
and the vorticity, which is zero in the incident current, remains practically
zero on these streamlines, as it would in an ideal fluid. Thus the flow at
great distances from the body may be regarded as potential flow everywhere
except in the wake.
We shall now derive formulae relating the properties of the flow in the
wake to the forces acting on the body. The total momentum transported by
the fluid through any closed surface surrounding the body is equal to the
t In contradistinction to the turbulent wake; see §36.
72 Viscous Fluids §21
integral of the momentum flux density tensor over that surface, <j> H adfjc.
The components of the tensor II $& are
n« = p&ik+ p(Ui+Vi)(Uic+v k ).
We write the pressure in the form p = po +p\ where po is the pressure at
infinity. The integration of the constant term poSac+ pUtUjc gives zero,
since the vector integral  df over a closed surface is zero. The integral
Ui <j> pvjcdfk also vanishes : since the total mass of fluid in the volume con
sidered is constant, the total mass flux <j> pv«df through the surface surround
ing the volume must be zero. Finally, the velocity v far from the body is
small compared with U. Hence, if the surface in question is sufficiently far
from the body, we can neglect the term pvwk in 11^ as compared with
pUjcVi. Thus the total momentum flux is
<j> (p'Sik + pUjcViWk
Let us now take the fluid volume concerned to be the volume between two
infinite planes x = constant, one of them far in front of the body and the
other far behind it. The integral over the infinitely distant "lateral" surface
vanishes (since p' — v = at infinity), and it is therefore sufficient to inte
grate only over the two planes. The momentum flux thus obtained is
evidently the difference between the total momentum flux entering through
the forward plane and that leaving through the backward plane. This
difference, however, is just the quantity of momentum transmitted to the body
by the fluid per unit time, i.e. the force F exerted on the body.
Thus the components of the force F are
^ = ( J7  \\)(p'+ P Uv x )dydz,
F ^ = ( J7  j S )pUvydydz >
F* = ( j f ~ j j )pUv z dydz,
where the integration is taken over the infinite planes x = x± (far behind the
body) and x = X2 (far in front of it). Let us first consider the expression for
F x .
Outside the wake we have potential flow, and therefore Bernoulli's equation
p+%p(U+v) 2 = constant = p + %pU 2
holds, or, neglecting the term \pv 2 in comparison with />U«v,
p' =  P Uv x .
JJ«*d>.d«J'jgd V d,0, Jjgd,d»0,
§21 The laminar wake 73
We see that in this approximation the integrand in F x vanishes everywhere
outside the wake. In other words, the integral over the plane x = #2 (which
lies in front of the body and does not intersect the wake) is zero, and the
integral over the plane x = xi need be taken only over the area covered by
the crosssection of the wake. Inside the wake, however, the pressure change
p' is of the order of pv 2 , i.e. small compared with pUv x . Thus we reach the
result that the drag on the body is
F x =  P UJjv x dydz, (21.1)
where the integration is taken over the crosssectional area of the wake far
behind the body. The velocity v x in the wake is, of course, negative: the
fluid moves more slowly than it would if the body were absent. Attention is
called to the fact that the integral in (21.1) gives the amount by which the
discharge through the wake falls short of its value in the absence of the body.
Let us now consider the force (whose components are F y , F z ) which tends
to move the body transversely. This force is called the lift. Outside the
wake, where we have potential flow, we can write v y = 8<f>jdy, v z = dc/>J8z;
the integral over the plane x = X2, which does not meet the wake, is zero:
8< f> , /x f f d( f>
dy
since <j> = at infinity. We therefore find for the lift
F y = pUJj v y dy dz, F z =  P U jj v z dy dz. (21.2)
The integration in these formulae is again taken only over the crosssectional
area of the wake. If the body has an axis of symmetry (not necessarily
complete axial symmetry), and the flow is parallel to this axis, then the flow
past the body has an axis of symmetry also. In this case the lift is, of course,
zero.
Let us return to the flow in the wake. An estimate of the magnitudes of
various terms in the NavierStokes equation shows that the term »>A v can
in general be neglected at distances r from the body such that rUjv > 1
(cf . the derivation of the opposite condition at the beginning of §20) ; these
are the distances at which the flow outside the wake may be regarded as
potential flow. It is not possible to neglect that term inside the wake even
at these distances, however, since the transverse derivatives 8 2 vfdy 2 , dfyjdz 2
are large compared with dfy/dx 2 .
The term (v«grad)v in the NavierStokes equation is of the order of mag
nitude (U+v)dvjdx ~ Uvjx in the wake. The term vAv is of the order
of vd 2 v[dy 2 ~ w/Y 2 , where Y denotes the width of the wake, i.e. the order
of magnitude of the distances from the #axis at which the velocity v falls
off markedly. If these two magnitudes are comparable, we find
Y ~ ^(vxlU). (21.3)
74 Viscous Fluids §21
This quantity is in fact small compared with x, by the assumed condition
Uxjv > 1. Thus the width of the laminar wake increases as the square root
of the distance from the body.
In order to determine how the velocity decreases with increasing x in the
wake, we return to formula (21.1). The region of integration has an area of
the order of Y 2 . Hence the integral can be estimated as F x ~ pUvY 2 ,
and by using the relation (21.3) we obtain
v ~ Fzlpvx. (21.4)
PROBLEMS
Problem 1. Determine the flow in the laminar wake when there is both drag and lift.
Solution. Writing the velocity in the NavierStokes equation in the form U+v and
omitting terms quadratic in v (far from the body) we obtain
— = gra y\+ v y— + —y y
df
we have also neglected the term 8 2 v/8x 2 in Av. We seek a solution in the form v = v 1 +v a>
where v x satisfies
dvi /d 2 vi 8 2 vi\
U — = v + .
dx \ dy 2 dz 2 )
The term v 2 , which appears because of the term — grad(p/p) in the original equation, may be
taken as the gradient grad O of some scalar. Since the derivatives with respect to x, far from
the body, are small in comparison with those with respect to y and z, we may to the same
approximation neglect the term 8Q>ldx in v x , i.e. take v x = v ix .
Thus we have for v x the equation
dv x / d*v x d*v x \
U = v  + r .
dx \ dy 2 dz 2 J
This equation is formally the same as the twodimensional equation of heat conduction, with
x/U in place of the time, and the viscosity v in place of the thermometric conductivity.
The solution which decreases with increasing y and z (for fixed *) and gives an infinitely
narrow wake as * »■ (in this approximation the dimensions of the body are regarded as
small) is (see §51)
Fr 1
Vx= £_ _ e l7<y , +* , >/4r* (1)
Artpv X
The constant coefficient in this formula is expressed in terms of the drag by means of
formula (21.1), in which the integration over y and z may be extended to ±oo on account
of the rapid decrease of v x . If we replace the Cartesian coordinates by spherical coordinates
r, 9, <f> with the polar axis along the *axis, then the region of the wake (V(y 2 + J8 *) <^*)
corresponds to 9 <^ 1. In these coordinates formula (1) becomes
ATrpv r
The term dQ>/dx (with O given by formula (3) below), which we have omitted, would give
a term in v x which diminishes more rapidly, as 1/r 2 .
§21 The laminar wake 75
v iv and v lz must have the same form as (1). We take the direction of the lift as the yaxis
(so that F t = 0). According to (21.2) we have, since O — at infinity,
oo oo
J J Myd«JJ («* + —)*>&
—oo —oo
= jj viy dydz =  FyjpU,
jjvudydz = 0.
Determining the constants in v iy and v\ t from these conditions, we find
F v i , , ao 8®
v y = e m^Vtox + 1 VgSS — . (2)
Airpv x By Bz
To determine the function O we proceed as follows. By the equation of continuity,
div v « dvy/dy + dvzldz = 0;
substituting (2), we have
\ dy* dz 2 / By
Differentiating this equation with respect to * and using the equation satisfied by v ly , we
obtain
/a 2 8 2 \8® a / dviy \
\ By 2 Bz 2 ) Bx ~ By\ Bx J
(■
Hence
a 2 a 2 \ dviy
+ — \ .
ay 2 Bz 2 1 By
BQ> v Bviy
Bx U By
Finally, substituting the expression for v iy and integrating with respect to x, we have
F v V
<I> = " Teutf+zv&x 1], (3)
27TpU y 2 + z 2
The constant of integration is chosen so that G> remains finite when y = z = 0. In spherical
coordinates (with the azimuthal angle <f> measured from the xyplane)
F v cos 6
It is seen from (2) and (3) that v y and v z , unlike v x , contain terms which decrease only as 1/0 2
as we move away from the "axis" of the wake, as well as those which decrease exponen
tially with 6 (for a given r).
The qualitative results (21.3) and (21.4) are, as we should expect, in agreement with the
above formulae. If there is no lift, the flow in the wake is axially symmetrical.
76 Viscous Fluids §22
Problem 2. Determine the flow outside the wake far from the body.
Solution. Outside the wake we assume potential flow. Since we are interested only in
the terms in the potential <E> which decrease least rapidly with distance, we seek a solution of
Laplace's equation A$ = as a sum of two terms:
a cos<£
r r
of which the first is centrally symmetric and belongs to the force F x , while the second is
symmetrical about the xyplane and belongs to the force F v .
Using the expression for A® in spherical coordinates, we obtain for the function
/(#) the equation
— (sin0— )
dd\ del
J » _£_ = o.
sin 6
The solution of this equation finite as 6 > n is/ = b cot £0. The coefficient b must be deter
mined so as to give the correct value of F y . It is simpler, however, to use the fact that in the
range V( v /Ur) <^ 8 <^ 1 this part of <S> must be the same as the expression
o = Fy C0S ^
TmpU r6 '
obtained from formula (3'), Problem 1, for O in the wake. Hence b = FJAnpU.
To determine the coefficient a, we notice that the total mass flux through a sphere S of
large radius r equals zero, as for any closed surface. The rate of inflow through the part S
of 5 intercepted by the wake is
 JJvxdydz = F x jpU.
S
Hence the same quantity must flow out through the rest of the surface of the sphere, i.e. we
must have
<f vdf= F X / P U.
ss
Since S is small compared with S, we can put
j>vd{ = ^gradOdf = Ana = F x j P U,
s s
whence a = —FxftnpU.
The complete solution is given by the sum of these two expressions :
1
which gives the flow everywhere outside the wake far from the body. The potential dimini
shes with increasing distance as 1/r; the velocity v, therefore, diminishes as 1/r 2 . If there is
no lift, the flow outside the wake is spherically symmetrical.
§22. The viscosity of suspensions
A fluid in which numerous fine solid particles are suspended (forming a
§22 The viscosity of suspensions 77
suspension) may be regarded as an homogeneous medium if we are concerned
with phenomena whose characteristic lengths are large compared with the
dimensions of the particles. Such a medium has an effective viscosity r\
which is different from the viscosity rjo of the original fluid. The value of rj
can be calculated for the case where the concentration of the suspended
particles is small (i.e. their total volume is small in comparison with that
of the fluid). The calculations are relatively simple for the case of spherical
particles (A. Einstein, 1906).
It is necessary to consider first the effect of a single solid globule, immersed
in a fluid, on flow having a constant velocity gradient. Let the unperturbed
flow be described by a linear velocity distribution
vot = aikX/c, (22.1)
where a^ is a constant symmetrical tensor. The fluid pressure is constant:
Po = constant,
and in future we shall take po to be zero, i.e. measure only the deviation
from this constant value. If the fluid is incompressible (div vo = 0), the
sum of the diagonal elements of the tensor a^ must be zero :
am = 0. (22.2)
Now let a small sphere of radius R be placed at the origin. We denote
the altered fluid velocity by v = vo + Vi ; vi must vanish at infinity, but near
the sphere vi is not small compared with Vo. It is clear from the symmetry
of the flow that the sphere remains at rest, so that the boundary condition is
v = for r = R.
The required solution of the equations of motion (20.1) to (20.3) may be
obtained at once from the solution (20.4), with the function /given by (20.6),
if we notice that the space derivatives of this solution are themselves solutions.
In the present case we desire a solution depending on the components of the
tensor a^ as parameters (and not on the vector u as in §20). Such a solution
is
vi = curl curl [(ocgrad)/], p = rjoCLwd 2 Afjdxidxn,
where (a »grad)/ denotes a vector whose components are oLnfifjdxjc. Expand
ing these expressions and determining the constants a and b in the function
/ = ar + bjr so as to satisfy the boundary conditions at the surface of the
sphere, we obtain the following formulae for the velocity and pressure :
5/RS R*\ #5
vu =  — WkVUnm ocm;%, (22.3)
2 \ r 4 r l l r 4
p =  5170—  (xikninjc, (22 A)
r 3
where n is a unit vector in the direction of the radius vector.
78 Viscous Fluids §22
Returning now to the problem of determining the effective viscosity of a
suspension, we calculate the mean value (over the volume) of the momentum
flux density tensor 11^, which, in the linear approximation with respect
to the velocity, is the same as the stress tensor — a tt :
*» = (1/V) j <j ik dV.
The integration here may be taken over the volume V of a sphere of large
radius, which is then extended to infinity.
First of all, we have the identity
/ dv t dv k \
°ik = yo — + —  I pS ik +
\ OXjc OXi I
f /(•**(£ + 5) ^K (22  5)
+
The integrand on the right is zero except within the solid spheres; since
the concentration of the suspension is supposed small, the integral may be
calculated for a single sphere as if the others were absent, and then multiplied
by the concentration c of the suspension (the number of spheres per unit
volume). The direct calculation of this integral would require an investi
gation of internal stresses in the spheres. We can circumvent this difficulty,
however, by transforming the volume integral into a surface integral over an
infinitely distant sphere, which lies entirely in the fluid. To do so, we note
that the equation of motion daujdxi = leads to the identity
<*ik = 8(cruXk)ldxi;
hence the transformation of the volume integral into a surface integral gives
<*ik
I dvi dv k \ r
= ^o(— + — l+cd) {auXkdfiTjo^idfk + vjcdfi)}.
We have omitted the term in p, since the mean pressure is necessarily zero ;
p is a scalar, which must be given by a linear combination of the components
onijc, and the only such scalar is ecu = 0.
In calculating the integral over a sphere of very large radius, only the
terms of order 1/r 2 need be retained in the expression (22.3) for the velocity.
A simple calculation gives the value of the integral as
cr]o • 207rR s {5a.i m ninjcnin m —aan]cni},
where the bar denotes an average with respect to directions of the unit vector
§23 Exact solutions of the equations of motion for a viscous fluid 79
n. Effecting the averaging, j we finally have
(dVi dvjc \
1 I + 577o«» '%ttR z c. (22.6)
oxjc 8xi J
The ratio of the second term to the first determines the required relative
correction to give the effective viscosity of the suspension. If we are in
terested only in corrections of the first order of smallness, we can take the
first term as 2^oa«Ar We then obtain for the effective viscosity of the suspen
sion
r) = r}0(l+m> (22.7)
where <f> = &nR z c is the small ratio of the total volume of the spheres to
the total volume of the suspension.
§23. Exact solutions of the equations of motion for a viscous fluid
If the nonlinear terms in the equations of motion of a viscous fluid do
not vanish identically, the solving of these equations offers great difficulties,
and exact solutions can be obtained only in a very small number of cases.
Furthermore, it has not yet proved possible to carry out a complete investi
gation of the steady flow of a viscous fluid in all space round a body in the
limit of very large Reynolds numbers. Although, as we shall see, such
a flow does not in practice remain steady, the solution of the problem would
nevertheless be of great methodological interest. $
We give below examples of exact solutions of the equations of motion for
a viscous fluid.
(1) An infinite plane disk immersed in a viscous fluid rotates uniformly
about its axis. Determine the motion of the fluid caused by this motion of
the disk (T. von KArman, 1921).
We take cylindrical coordinates, with the plane of the disk as the plane
z = 0. Let the disk rotate about the #axis with angular velocity Q. We
consider the unbounded volume of fluid on the side z > 0. The boundary
conditions are
*V = 0,
»* = Qr,
v z = for z = 0,
v r = 0,
^ =
for z = oo.
t The required mean values of products of components of the unit vector are symmetrical tensors,
which can be formed only from the unit tensor S,*. We then easily find
ntfljc = hoc,
ninicmn m = Tt(8ik$lm + 8il$km + $im$ki)
X The "vanishing viscosity" theory of Oseen is concerned with this problem; it is unsatis
factory, since it is based on an unjustified simplification of the NavierStokes equations. Prandtl's
boundarylayer theory (see §39) does not solve the problem throughout the volume of the fluid.
80
Viscous Fluids
§23
The axial velocity v z does not vanish as z > oo, but tends to a constant nega
tive value determined by the equations of motion. The reason is that,
since the fluid moves radially away from the axis of rotation, especially near
the disk, there must be a constant vertical flow from infinity in order to
satisfy the equation of continuity. We seek a solution of the equations of
motion in the form
(23.1)
v r = rQF(z!); ^ = rQG(*i); v z = i/(yQ)H(xi);
p — —pv£lP(zi) y where z\ = \/(Q/v)z.
In this velocity distribution, the radial and azimuthal velocities are propor
tional to the distance from the axis of rotation, while v z is constant on each
horizontal plane.
Substituting in the NavierStokes equation and in the equation of con
tinuity, we obtain the following equations for the functions F, G, H and P:
F2_ G*+F'H = F", 2FG+ G'H = G",
(23 2^
HH' = P'+H", 2F+H' = 0; V ' '
the prime denotes differentiation with respect to z\. The boundary conditions
are
F = 0, G = 1, H = for *i = 0.
(23.3)
F = 0, G = for ^i = oo. v }
We have therefore reduced the solution of the problem to the integration of a
system of ordinary differential equations in one variable ; this can be achieved
numerically.f Fig. 7 shows the functions F, G and — H thus obtained.
t The numerical integration has also been carried out for another similar problem, in which the
fluid rotates uniformly at infinity and the disc is at rest (U.T. Bodewadt, Zeitschrift fur angeviandte
Mathematik und Mechanik 20, 241, 1940).
§23 Exact solutions of the equations of motion for a viscous fluid 81
The limiting value of H as zi > oo is 0886; in other words, the fluid
velocity at infinity is fl z (oo) = — Q886\/(vQ).
The frictional force acting on unit area of the disk perpendicularly to the
radius is a z<f) = rj{dv^dz) z= Q. Neglecting edge effects, we may write the
moment of the frictional forces acting on a disk of large but finite radius R as
R
M = 2 j 2rrr 2 a H Ar = 7r#W0^ 3 )G'( )
o
The factor 2 in front of the integral appears because the disk has two sides
exposed to the fluid. A numerical calculation of the function G leads to
the formula
M =  194 JRW("Q 8 ) ( 23  4 )
(2) Determine the steady flow between two plane walls meeting at an
angle a (Fig. 8 shows a crosssection of the two planes); the fluid flows
out from the line of intersection of the planes (G. Hamel, 1916).
Fig. 8
We take cylindrical coordinates r, z, <f>, with the zaxis along the line of
the intersection of the planes (the point O in Fig. 8), and the angle <£ measured
as shown in Fig. 8. The flow is uniform in the ^direction, and we naturally
assume it to be entirely radial, i.e.
i>4 = Vz = 0, v r = v(r, <f>).
The equations (15.16) give
dv 1 dp I d 2 v 1 d 2 v 1 dv v \
*,_.= L + J — + + , (23.5)
8r pdr \ dr 2 r d<f> r dr r 2 /
1 dp 2v dv ,_ , x
£ + ^^r ' (23  6)
pr d<f> r* d<j>
d(rv)jdr = 0.
It is seen from the last of these that rv is a function of <f> only. Introducing
the function
u(<f>) = rvl6v, (23.7)
82
Viscous Fluids
we obtain from (23.6)
1 dp \2v* du
pd<j>~ r2 d<j>'
whence
p 12l/ 2
§23
ft
Substituting this expression in (23.5), we have
dhi 1
_ + 4 „ + 6«s = —,*/<(,),
from which we see that, since the lefthand side depends only on <£ and the
righthand side only on r, each must be a constant, which we denote by 2C\.
Thus /'(*■) = 12v 2 Ci/r 3 , whence /(r) = 6v 2 Ci/r 2 + constant, and we have
for the pressure
p 6v 2
 = —r(2u  d) + constant. (23 .8)
p r z
For u(<f>) we have the equation
m" + 4m + 6« 2 = 2Ci,
which, on multiplication by u' and one integration, gives
m' 2 + 2 m 2 + 2 w 3_2Cim2C 2 = 0.
Hence we have
24 = ± + C 3 , (23.9)
which gives the required dependence of the velocity on <f>; the function u(<f>)
can be expressed in terms of elliptic functions. The three constants Ci, C2,
Cz are determined from the boundary conditions
«(±i<x) = (23.10)
and from the condition that the same mass Q of fluid passes in unit time
through any crosssection r = constant:
a/2 a/2
J vrd<f> = 6vp J
a/2 a/2
Q = p j vrdt = 6v P j ud<f>. (23.11)
Q may be either positive or negative. If Q > 0, the line of intersection
of the planes is a source, i.e. the fluid emerges from the vertex of the angle :
this is called flow in a diverging channel. If Q < 0, the line of intersection is
§23 Exact solutions of the equations of motion for a viscous fluid 83
a sink, and we have flow in a converging channel. The ratio \Q\jvp is dimen
sionless and plays the part of the Reynolds number in the problem considered.
Let us first discuss converging flow (Q < 0). To investigate the solution
(23.9)(23.11) we make the assumptions, which will be justified later, that
the flow is symmetrical about the plane <f> = (i.e. u(<f>) = u( — <f>)), and that
the function u(4>) is everywhere negative (i.e. the velocity is everywhere
towards the vertex) and decreases monotonically from u = at <j> = ± fa
to u = — «o < at 4> = 0, so that uq is the maximum value of \u\. Then
for u = —wo we must have dufd</> = 0, whence it follows that u = — #o
is a zero of the cubic expression under the radical in the integrand of (23.9).
We can therefore write
— u z —u 2 +Ciu+C2 = (u + uo){ — u 2 (luo)u + q} f
where q is another constant. Thus
u
C dw
26 = + , (23.12)
^ "J V\(u + u ){u*auo)u + q\y
V[( M + «o){  m 2  (1  mo)w + q}] '
the constants uq and q being determined from the conditions
o
a= f
«,
du
J VKu
V[(« + «o){  m 2  (1  u )u + q}] '
(23.13)
udu
\/[(u + u ){  m 2  (1  u )u + q}]
1*0
Fig. 9
(R = \Q\jvp)\ the constant q must be positive, since otherwise these integrals
would be complex. The two equations just given may be shown to have
solutions uq and q for any R and a < it. In other words, convergent sym
metrical flow (Fig. 9) is possible for any aperture angle a and any Reynolds
84 Viscous Fluids §23
number. Let us consider in more detail the flow for very large R. This
corresponds to large uq. Writing (23.12) (for <j> > 0) as
j
r d«
2(a<£) = ,
J \Z[(u + uo){u 2 (lu )u + q}]
we see that the integrand is small throughout the range of integration if \u \
is not close to uq. This means that \u  can differ appreciably from uq only
for <f> close to ±^a, i.e. in the immediate neighbourhood of the walls.f
In other words, we have u « constant = — uq for almost all angles <f>, and
in addition uq = R/6a, as we see from equations (23.13). The velocity v
itself is \Q /potr, giving a nonviscous potential flow with velocity independent
of angle and inversely proportional to r. Thus, for large Reynolds numbers,
the flow in a converging channel differs very little from potential flow of
an ideal fluid. The effect of the viscosity appears only in a very narrow layer
near the walls, where the velocity falls rapidly to zero from the value cor
responding to the potential flow (Fig. 10).
Fig. 10
Now let Q > 0, so that we have divergent flow. At first we again suppose
that the flow is symmetrical about the plane <j> = 0, and that u{<j>) (where
now u > 0) varies monotonically from zero at cf> = ± a to uq > at <j> = 0.
Instead of (23.13) we now have
■/
«0 ,
QU
V[(«o  u){u 2 + (1 + uq)u + q}] '
(23.14)
r udu
J vT( M o — u){u 2 + (l + uo)u + q}]
f The question may be asked how the integral can cease to be small, even if u 7H — u . The answer
is that, for u very large, one of the roots of — w 2 —(1 — w )u+g = is close to — u , so that the radicand
has two almost coincident zeros, the whole integral therefore being "almost divergent" at u = —u .
§23 Exact solutions of the equations of motion for a viscous fluid 85
If we regard u as given, then a increases monotonically as q decreases, and
takes its greatest value for q = 0:
du
CCmax
r du
J a/\u(uou)(u
\/[u(uq  u)(u + UQ + 1)]
Fig. 11
It is easy to see that for given q, on the other hand, a is a monotonically
decreasing function of uq. Hence it follows that uq is a monotonically de
creasing function of q for given a, so that its greatest value is for q =
and is given by the above equation. The maximum R = Rmax corresponds
to the maximum wo Using the substitutions k 2 = uol(l + 2uo), u = uq cos 2 x,
we can write the dependence of R ma x on a in the parametric form
ax
C QX
a = 2V(l2tf) — — — — ,
J y\l— k z sm*x)
1&2 12
Rmax = — 6a — — +
tt/2
(23.15)
12*2 V(l2£ 2 )
\/(l — k 2 sin 2 x)dx.
Thus symmetrical flow, everywhere divergent (Fig. 11a), is possible for a
given aperture angle only for Reynolds numbers not exceeding a definite
value. As a > tr (k > 0), Rmax > 0; as a > (k > l/\/2)> Rmax tends to
infinity as 188/a.
For R > Rmax the assumption of symmetrical flow, everywhere divergent,
is unjustified, since the conditions (23.14) cannot be satisfied. In the range
of angles — a < <f> ^ \on the function u(<f>) must now have maxima or
minima. The values of u(<f>) corresponding to these extrema must again be
zeros of the polynomial under the radical sign. It is therefore clear that
the trinomial u 2 + (l+Uo)u + q (with uq > 0, q > 0) must have two real
negative roots in the range mentioned, so that the radicand can be written
{uq — u)(u+uo')(u+uo"), where uq > 0, uq' > 0, uq" > 0; we suppose
86 Viscous Fluids §23
wo' < «o". The function u{<j>) can evidently vary in the range mo > u ^ — uq\
u = uo corresponding to a positive maximum of u(<f>), and u — «o' to a
negative minimum. Without pausing to make a detailed investigation of the
solutions obtained in this way, we may mention that for R > R m ax a solution
appears in which the velocity has one maximum and one minimum, the flow
being asymmetric about the plane <f> = (Fig. lib). When R increases fur
ther, a symmetrical solution with one minimum and two maxima appears
(Fig. lie), and so on. In all these solutions, therefore, there are regions of
both outward and inward flow (though of course the total discharge Q
is positive). As R > oo the number of alternating minima and maxima
increases without limit, so that there is no definite limiting solution. We
may emphasise that in divergent flow as R> oo the solution does not,
therefore, tend to the solution of Euler's equations as it does for convergent
flow. Finally, it may be mentioned that, as R increases, the steady divergent
flow of the kind described becomes unstable soon after R exceeds R m ax>
and in practice a nonsteady or turbulent flow occurs (Chapter III).
(3) Determine the flow in a jet emerging from the end of a narrow tube
into an infinite space filled with the fluid — the submerged jet (L. Landau,
1943).
We take spherical coordinates r, 6, <f>, with the polar axis in the direction
of the jet at its point of emergence, and with this point as origin. The flow is
symmetrical about the polar axis, so that v^ = and v dl v r are functions of r
and 6 only. The same total momentum flux (the "momentum of the jet")
must pass through any closed surface surrounding the origin (in particular,
through an infinitely distant surface). For this to be so, the velocity must be
inversely proportional to r, so that
v r = F(d)fr, v e = f(d)[r, (23.16)
where F and / are some functions of 6 only. The equation of continuity is
1 d{r*v r ) 1 d
+ — (v g sin 6) = 0.
r 2 dr r sin 6 88
Hence we find that
F(d) = dflddfcot9. (23.17)
The components II r ^, 11^ of the momentum flux density tensor in the jet
vanish identically by symmetry. We assume that the components 11^
and 11^ also vanish; this assumption is justified when we obtain a solution
satisfying all the necessary conditions. Using the expressions (15.17) for
the components of the tensor o%, and formulae (23.16), (23.17), we easily
see that the relation
sin2 lire =  ^[sin2 0(II^rU)]
holds between the components of the momentum flux density tensor in the
§23 Exact solutions of the equations of motion for a viscous fluid 87
jet. Hence it follows that U r e = 0. Thus only the component II rr is non
zero, and it varies as 1/r 2 . It is easy to see that the equations of motion
dUijcjdXk = are automatically satisfied.
Next, we write
(n*n*)/p = (p+2vfcote2vf)ir2 = o,
or
d(l//)/d0+(l//)cot0+l/2i> = 0.
The solution of this equation is
/= 2i/sin0/(^cos0), (23.18)
and then we have from (23.17)
[ A 2  1 \
F = 2v 1 . (23.19)
l(,4cos0)2 / v '
The pressure distribution is found from the equation
IWP = Plp+f(f+2vcote)lr* = 0,
which gives
4pv 2 (^cos0l)
r z (A — cos 0) 2
The constant A can be found in terms of the momentum of the jet, i.e. the
total momentum flux in it. This flux is equal to the integral over the surface
of a sphere
n
P = <J> n„. cosfld/ = 2tt f r^Urr cos0sin0 d0.
The value of II rr is given by
( (A*l)* A \
{ (Acosd)* ~ Acosdy
1 4^2 ( (A 2  1) 2 A
p r 2 l(^4cos0)4 ^4cos0f
and a calculation of the integral gives
P =16 ^( 1+ _J__^ log ^±ij. (23 . 21)
Formulae (23.16)(23.21) give the solution of the problem.f
t The solution here obtained is exact for a jet regarded as emerging from a point source. If the
finite dimensions of the tube mouth are taken into account, the solution becomes the first term of an
expansion in powers of the ratio of these dimensions to the distance r from the mouth of the tube.
This is why, if we calculate from the above solution the total mass flux through a closed surface sur
rounding the origin, the result is zero. A nonzero total mass flux is obtained when further teuns
in the abovementioned expansion are considered; see Yu. B. Rumer, Prikladnaya matematika i
mekhanika 16, 255, 1952.
The submerged laminar jet with a nonzero angular momentum has bee ndiscussed by L. G.
LoItsyanskiI (ibid. 17, 3, 1953).
Viscous Fluids
§24
The streamlines are determined by the equation drfv r = rd0/v e , integration
of which gives r sin 2 6j{A — cos 6) = constant. Fig. 12 shows the streamlines
in the jet (for A > 1).
Fig. 12
Let us consider two limiting cases, a weak jet (small momentum P) and a
strong jet (large P). As P > 0, the constant A tends to infinity: from (23.21)
we have P = l&nv 2 pjA. For the velocity in this case we have
v d = —P sin 6 j&TTvpr, v r = P cos QjAm>pr.
As P > oo (strong jetf), A tends to unity: (23.21) gives A — l + a 2 , where
a = 32ttv 2 p/3P. For large angles (0 ~ 1), the velocity is given by
v 6 = — (2v/r) cotffl, v r = —2v/r,
but for small angles (0 ~ a) we have
^ = _4^/( a 2 + ^8), » r = 8va 2 /(a 2 +0 2 ) 2 .
§24. Oscillatory motion in a viscous fluid
When a solid body immersed in a viscous fluid oscillates, the flow thereby
set up has a number of characteristic properties. In order to study these,
it is convenient to begin with a simple but typical example. Let us suppose
that an incompressible fluid is bounded by an infinite plane surface which
executes a simple harmonic oscillation in its own plane, with frequency w.
We require the resulting motion of the fluid. We take the solid surface as the
yziplane, and the fluid region as x > ; the jyaxis is taken in the direction
of the oscillation. The velocity u of the oscillating surface is a function of
time, of the form A cos (atf + <x). It is convenient to write this as the real
part of a complex quantity:
u = re(uoe~ ia)t ),
where the constant uq = Ae i0L is in general complex, but can always be made
real by a proper choice of the origin of time.
t However, it must be borne in mind that the flow in a sufficiently strong jet is actually turbulent
(§35).
§24 Oscillatory motion in a viscous fluid 89
So long as the calculations involve only linear operations on the velocity
k, we may omit the sign re and proceed as if u were complex, taking the real
part of the final result. Thus we write
Uy = u = uo e"K (24.1)
The fluid velocity must satisfy the boundary condition v = u for x = 0,
i.e. v x = v 2 = 0, v y = u.
It is evident from symmetry that all quantities will depend only on the
coordinate x and the time t . From the equation of continuity div v =
we therefore have dv x /8x = 0, whence v x = constant = zero, from the
boundary condition. Since all quantities are independent of the coordinates
y and z, we have (v«grad)v = v x dyjdx, and since v x is zero it follows that
(v«grad)v = identically. The equation of motion (15.7) becomes
dv/dt = (lj P )gradp + vAv. (24.2)
This is a linear equation. Its ^component is dpjdx = 0, i.e. p = constant.
It is further evident from symmetry that the velocity v is everywhere in
the ^direction. For v y = v we have by (24.2)
dvjdt = vd^vjdx 2 , (24.3)
that is, a (onedimensional) heat conduction equation. We shall look for a
solution of this equation which is periodic in x and t, of the form
with a complex amplitude Mo, so that v = u for x = 0. Substituting in
(24.3), we find toy = vk 2 , whence
k = V( ioi l v ) = ± (*'+ WW 2 *)*
so that the velocity v is
v _ u e V(M/2v)z e iW(u/2v)xo)f} (24.4)
we have taken k to have a positive imaginary part, since otherwise the velocity
would increase without limit in the interior of the fluid, which is physically
impossible.
The solution obtained represents a transverse wave: its velocity v y — v
is perpendicular to the direction of propagation. The most important pro
perty of this wave is that it is rapidly damped in the interior of the fluid :
the amplitude decreases exponentially as the distance x from the solid
surface increases. f
Thus transverse waves can occur in a viscous fluid, but they are rapidly
damped as we move away from the solid surface whose motion generates the
waves.
The distance S over which the amplitude falls off by a factor of e is called
the depth of penetration of the wave. We see from (24.4) that
S = V(2v/w). (24.5)
t Over a distance of one wavelength the amplitude diminishes by a factor of c 2 "  K 540.
90 Viscous Fluids §24
The depth of penetration therefore diminishes with increasing frequency, but
increases with the kinematic viscosity of the fluid.
Let us calculate the frictional force acting on unit area of the plane oscil
lating in the viscous fluid. This force is evidently in the ^direction, and is
equal to the component a xy = v\dv y \dx of the stress tensor; the value of the
derivative must be taken at the surface itself, i.e. at x = 0. Substituting
(24.4), we obtain
°xy = Vd^pX* !)«• (24.6)
Supposing Mo real and taking the real part of (24.6), we have
<*xy = —\/(co7]p)uoCos(cot + l7r).
The velocity of the oscillating surface, however, is u = uq cos cot. There
is therefore a phase difference between the velocity and the frictional force.f
It is easy to calculate also the (time) average of the energy dissipation
in the above problem. This may be done by means of the general formula
(16.3); in this particular case, however, it is simpler to calculate the required
dissipation directly as the work done by the frictional forces. The energy
dissipated per unit time per unit area of the oscillating plane is equal to the
mean value of the product of the force a xy and the velocity u y = u:
 o zy u = i«o 2 \/(£ OM ?/>) (24.7)
It is proportional to the square root of the frequency of the oscillations,
and to the square root of the viscosity.
An explicit solution can also be given of the problem of a fluid set in
motion by a plane surface moving in its plane according to any law u = u(t).
We shall not pause to give the corresponding calculations here, since the
required solution of equation (24.3) is formally identical with that of an
analogous problem in the theory of thermal conduction, which we shall
discuss in §52 (the solution is formula (52.15)). In particular, the frictional
force on unit area of the surface is given by
°xy
yrjp r du(r) dr
7J^^) : (24 ' 8)
cf. (52.16).
t For oscillations of a half plane (parallel to its edge) there is an additional frictional force due to
edge effects. The problem of the motion of a viscous fluid caused by oscillations of a halfplane, and
also the more general problem of the oscillations of a wedge of any angle, can be solved by a class of
solutions of the equation A/+& 2 / *= 0, used by A. Sommehfeld in the theory of diffraction by a wedge;
see, for instance, M. von Laue, Interferenz und Beugung elektromagnetischer Wellen (Interference
and diffraction of electromagnetic waves), Handbuch der Experimentalphysik 18, 333, Akademische
Verlagsgesellschaft, Leipzig 1928.
We give here, for reference, only one result: the increase in the frictional force on a halfplane,
arising from the edge effect, can be regarded as the result of increasing the area of the halfplane by
moving the edge a distance JS = \Z(vj2to).
§24 Oscillatory motion in a viscous fluid 91
Let us now consider the general case of an oscillating body of arbitrary
shape. In the case of an oscillating plane considered above, the term
(v»grad)v in the equation of motion of the fluid was identically zero. This
does not happen, of course, for a surface of arbitrary shape. We shall assume,
however, that this term is small in comparison with the other terms, so that
it may be neglected. The conditions necessary for this procedure to be valid
will be examined below.
We shall therefore begin, as before, from the linear equation (24.2).
We take the curl of both sides; the term curlgradp vanishes identically,
giving
2(curlv)/3* = vAcurlv, (24.9)
i.e. curl v satisfies a heat conduction equation. We have seen above, however,
that such an equation gives an exponential decrease of the quantity which
satisfies it. We can therefore say that the vorticity decreases towards the
interior of the fluid. In other words, the motion of the fluid caused by the
oscillations of the body is rotational in a certain layer round the body, while
at larger distances it rapidly changes to potential flow. The depth of penetra
tion of the rotational flow is of the order of 8 ~ ^(vjcj).
Two important limiting cases are possible here: the quantity 8 may be
either large or small compared with the dimension of the oscillating body.
Let / be the order of magnitude of this dimension. We first consider the case
8 > /; this implies that 1 2 oj <^ v. Besides this condition, we shall also suppose
that the Reynolds number is small. If a is the amplitude of the oscillations,
the velocity of the body is of the order of aco. The Reynolds number for the
motion in question is therefore cual/v. We therefore suppose that
Pa> <^ v, waljv < 1. (24.10)
This is the case of low frequencies of oscillation, which in turn means that
the velocity varies only slowly with time, and therefore that we can neglect
the derivative dvfdt in the general equation of motion. The term (v»grad)v,
on the other hand, can be neglected because the Reynolds number is small.
The absence of the term dv]dt from the equation of motion means that the
flow is steady. Thus, for 8 > /, the flow can be regarded as steady at any
given instant. This means that the flow at any given instant is what it would
be if the body were moving uniformly with its instantaneous velocity. If,
for example, we are considering the oscillations of a sphere immersed in the
fluid, with a frequency satisfying the inequalities (24.10) (/ being now the
radius of the sphere), then we can say that the drag on the sphere will be that
given by Stokes' formula (20.14) for uniform motion of the sphere at small
Reynolds numbers.
Let us now consider the opposite case, where / > 8. In order that the
term (v«grad)v should again be negligible, it is necessary that the amplitude
of the oscillations should be small in comparison with the dimensions of the
body:
Pco >v, a<$l; (24.11)
92 Viscous Fluids §24
in this case, it should be noticed, the Reynolds number need not be small.
The above inequality is obtained by estimating the magnitude of (v«grad)v.
The operator (v»grad) denotes differentiation in the direction of the velocity.
Near the surface of the body, however, the velocity is nearly tangential. In
the tangential direction the velocity changes appreciably only over distances
of the order of the dimension of the body. Hence
(vgrad)v ~ v 2 jl ~ a 2 a> 2 /l,
since the velocity itself is of the order of am. The derivative d\jdt, however,
is of the order of vco ~ aco 2 . Comparing these, we see that
(v^grad)v 4 d\\dt
if a < /. The terms dvjdt and v A v are then easily seen to be of the same
order.
We may now discuss the nature of the flow round an oscillating body when
the conditions (24.11) hold. In a thin layer near the surface of the body
the flow is rotational, but in the rest of the fluid we have potential flow.f
Hence the flow everywhere except in the layer adjoining the body is given
by the equations
curlv = 0, div v = 0. (24.12)
Hence it follows that Av = 0, and the NavierStokes equation reduces to
Euler's equation. The flow is therefore ideal everywhere except in the
surface layer. Since this layer is thin, in solving equations (24.12) to deter
mine the flow of the rest of the fluid we should take as boundary conditions
those which must be satisfied at the surface of the body, i.e. that the fluid
velocity is equal to that of the body. The solutions of the equations of motion
for an ideal fluid cannot satisfy these conditions, however. We can require
only the fulfilment of the corresponding condition for the fluid velocity
component normal to the surface.
Although equations (24.12) are inapplicable in the surface layer of fluid,
the velocity distribution obtained by solving them satisfies the necessary
boundary condition for the normal velocity component, and the actual
variation of this component near the surface therefore has no significant
properties. The tangential component would be found, by solving the equa
tions (24.12), to have some value different from the corresponding velocity
component of the body, whereas these velocity components should be equal
also. Hence the tangential velocity component must change rapidly in the
surface layer. The nature of this variation is easily determined. Let us
consider any portion of the surface of the body, of dimension large compared
f For oscillations of a plane surface not only curl v but also v itself decreases exponentially with
characteristic distance 8. This is because the oscillating plane does not displace the fluid, and there
fore the fluid remote from it remains at rest. For oscillations of bodies of other shapes the fluid is
displaced, and therefore executes a motion where the velocity decreases appreciably only over distances
of the order of the dimension of the body.
§24 Oscillatory motion in a viscous fluid 93
with S, but small compared with the dimension of the body. Such a portion
may be regarded as approximately plane, and therefore we can use the re
sults obtained above for a plane surface. Let the #axis be directed along
the normal to the portion considered, and the ^yaxis parallel to the tangential
velocity component of the surface there. We denote by v y the tangential
component of the fluid velocity relative to the body; v y must vanish on the
surface. Lastly, let voe~ ia)t be the value of v y found by solving equations
(24.12). From the results obtained at the beginning of this section, we can
say that in the surface layer the quantity v y will fall off towards the surface
according to the law
Vy = VQ et»4[ler0O*vW2r)] m (24.13)
Finally, the total amount of energy dissipated in unit time will be given by
the integral
ikin = W(blP°>) j> bo 2 d/ (24.14)
taken over the surface of the oscillating body.
In the Problems at the end of this section we calculate the drag on various
bodies oscillating in a viscous fluid. Here we shall make the following general
remark regarding these forces. Writing the velocity of the body in the complex
form u = uoe iwt , we obtain a drag F proportional to the velocity u, and also
complex: F = /?«, where f3= fii + ife is a complex constant. This expression
can be written as the sum of two terms with real coefficients :
F = (pi + ife)u = faufeujco, (24.15)
one proportional to the velocity u and the other to the acceleration u.
The (time) average of the energy dissipation is given by the mean product
of the drag and the velocity, where of course we must first take the real
parts of the expressions given above, i.e. u = \{uoe~ i<at + uo*e io>t ),
F = l{uofie iwt + uo*p*e i0it ). Noticing that the mean values of e ±Zia)t are
zero, we have
Fu = l(P+P>) H 2 = ifr M 2  (24.16)
Thus we see that the energy dissipation arises only from the real part of /?;
the corresponding part of the drag (24.15), proportional to the velocity, may
be called the dissipative part. The other part of the drag, proportional to
the acceleration and determined by the imaginary part of /?, does not involve
the dissipation of energy and may be called the inertial part.
Similar considerations hold for the moment of the forces on a body execut
ing rotary oscillations in a viscous fluid.
PROBLEMS
Problem 1. Determine the frictional force on each of two parallel solid planes, between
which is a layer of viscous fluid, when one of the planes oscillates in its own plane.
94 Viscous Fluids §24
Solution. We seek a solution of equation (24.3) in the formf
v = (A sinkx+B coskx)e~ i<ot ,
and determine A and B from the conditions v = u = M e~ i& " for * = and v — for
x — h, where h is the distance between the planes. The result is
sin k(h — x)
V = u
sinkh
The frictional force per unit area on the moving plane is
Plx = rj(dv/dx)x=0 = —rjkucotkh,
while that on the fixed plane is
P<l x = — 7}(8vldx) x =h = f\hi cosec kh,
the real parts of all quantities being understood.
Problem 2. Determine the frictional force on an oscillating plane covered by a layer of
fluid of thickness h, the upper surface being free.
Solution. The boundary condition at the solid plane is v = u for * ■= 0, and that at the
free surface is a xy = rjdv/dx = for x = h. We find the velocity
COS k(h — x)
V = u .
coskh
The frictional force is
P x = 7)(dvjdx) x =^o — yku tan kh.
Problem 3. A plane disk of large radius R executes rotary oscillations of small amplitude
about its axis, the angle of rotation being = 8 cos <at, where & <^ 1 . Determine the moment
of the frictional forces acting on the disk.
Solution. For oscillations of small amplitude the term (v.grad)v in the equation of
motion is always small compared with 8v/8t, whatever the frequency to. If R ^> 8, the disk
may be regarded as infinite in determining the velocity distribution. We take cylindrical
coordinates, with the jaraxis along the axis of rotation, and seek a solution such that
Vr = Vt = 0, v* = v = rQ(s, t). For the angular velocity Q(z, t) of the fluid we obtain the
equation
BQjdt = vdm/dz 2 .
The solution of this equation which is — w6 Q sin tot for z = and zero for z = oo is
Q =  <od e~ z/s sin(wt  z/8).
The moment of the frictional forces on both sides of the disk is
R
M = 2 f r'27rrr)(dvldz) s= >odr = wOvn^/(u)prq)R* cos(arfiir).
t In all the Problems to this section S denotes the quantity (24.5):
8 = V( 2v h)> and k = (l + i)IS.
§24 Oscillatory motion in a viscous fluid 95
Problem 4. Determine the flow between two parallel planes when there is a pressure
gradient which varies harmonically with time.
Solution. We take the warplane halfway between the two planes, with the waxis parallel
to the pressure gradient, which we write in the form
(l/p)dp/dx = ae~ i(at .
The velocity is everywhere in the ^direction, and is determined by the equation
dv/dt = aeM+vdHldy*.
The solution of this equation which satisfies the conditions v — for y = ±JA is
COS Ay
COS^M.
The mean value of the velocity over a crosssection is
2
ia r cos Ay 1
v = — e itot \ 1 — .
co I cosiMj
ia l I \
v = — e~ ia>t 1 tan AM .
co \ kh * /
For h/8 <^ 1 this becomes
v « aet^h 2 1 12v,
in agreement with (17.5), while for hjh ^> 1 we have
v x (ia/co)e~ i<at ,
in accordance with the fact that in this case the velocity must be almost constant over the
crosssection, varying only in a narrow surface layer.
Problem 5. Determine the drag on a sphere of radius R which executes translatory oscil
lations in a fluid.
Solution. We write the velocity of the sphere in the form u = u c _ * <u '. As in §20, we
seek the fluid velocity in the form v = e~ ib)t curl curl/u , where / is a function of r only
(the origin is taken at the instantaneous position of the centre of the sphere). Substituting in
(24.9) and effecting transformations similar to those of §20, we obtain the equation
A a /+(M>W A/=0
(instead of the equation A 2 / = in §20). Hence we have
A/ = constant x e ikr /r,
the solution being chosen which decreases exponentially with r. Integrating, we have
df[dr = [ae^r(rlfik) + b]lr^; (1)
the function/ itself is not needed, since only the derivatives/ ' and/ " appear in the velocity.
The constants a and b are determined from the condition that v = u for r = R, and are
found to be
3R / 3 3 \
a =  —ei**, b = W[\ . (2)
lik ¥ \ ikR k*R*I K)
It may be pointed out that, at large distances {R ^> 8), a > and b > — \R* t the values for
potential flow obtained in §10, Problem 2; this is in accordance with what was said in §24.
96 Viscous Fluids §24
The drag is calculated from formula (20.13), in which the integration is over the surface
of the sphere. The result is
/ R \ I 2R \ du
F = 6^1 +  )u + 3iTRW(2rip/co)(l + — J — . (3)
For to = this becomes Stokes' formula, while for large frequencies we have
du
F = %tt P R* + 2>TTR^(2r }P oi)u.
dt
The first term in this expression corresponds to the inertial force in potential flow past a
sphere (see §11, Problem 1), while the second gives the limit of the dissipative force.
Problem 6. Determine the drag on a sphere moving in an arbitrary manner, the velocity
being, given by a function u{t).
Solution. We represent u(t) as a Fourier integral :
f ! f
M (0 = u^e^tdco, u u = — u{r)e ib>T dr.
J 2tt J
— oo —oo
Since the equations are linear, the total drag may be written as the integral of the drag forces
for velocities which are the separate Fourier components u (t) e~ iwt ] these forces are given by
(3) of Problem 5, and are
Noticing that (du/dt)^ = —icou^, we can rewrite this as
( 6v 3V(2v) 1 + M
On integration over to, the first and second terms give respectively u(t) and u(t). To integrate
the third term, we notice first of all that for negative co this term must be written in the
complex conjugate form, (1 +i)j\/co being replaced by (1 — *)/ 'VMi tm s is because formula
(3) of Problem 5 was derived for a velocity u — u e~ i<Jt with to > 0, and for a velocity
«„£*«' we should obtain the complex conjugate. Instead of an integral over w from — oo
to + oo, we can therefore take twice the real part of the integral from to oo. We write
oo, . , . oo oo
2iel(l+i) K ^ da) = re (1 + ^ dwdr
oo
oo oo oo
1 i r r u(t) e ib *tr) r /• u (t) eft****
re (1+0 ^ dcodr + (l + ^ dcodi
IT \ J J \/CO J J S/Oi
oo *
t 00
Vtt \ J V(tr) J V(t0 J
N IT J
—oo
t
U(r)
V(tr\
dr.
§24 Oscillatory motion in a viscous fluid 97
Thus we have finally for the drag
i\du 3vu 3 iv rdu dr )
f=2lrpi? 3{__ + _ + _yj_ j. (1)
— oo
Problem 7. Determine the drag on a sphere which at time t = begins to move with a
uniform acceleration, u = at.
Solution. Putting, in formula (1) of Problem 6, u = for t < and u = at for t >
we have for t >
rl 3vt 6 /tvi
F = 2.pfi3«[ +  +  A /J.
Problem 8. The same as Problem 7, but for a sphere brought instantaneously into uniform
motion.
Solution. We have u = f or t < and u = u for t > 0. The derivative dujdt is
zero except at the instant t = 0, when it is infinite, but the time integral of du/dt is finite,
and equals « . As a result, we have for all t >
[D i
1 + —  —  +&rpR 3 U 8(t),
V(irrt) J
where B(t) is the delta function. For t > oo this expression tends asymptotically to the value
given by Stokes' formula. The impulsive drag on the sphere at t = is obtained by integrat
ing the last term and is %irpR 3 u .
Problem 9. Determine the moment of the forces on a sphere executing rotary oscillations
about a diameter in a viscous fluid.
Solution. For the same reasons as in §20, Problem 1, the pressuregradient term can be
omitted from the equation of motion, so that we have dv/dt = v A v. We seek a solution in
the form v = curl/£2 <r i6 ' t > where SI = Sl e~ ib)t is the angular velocity of rotation of the
sphere. We then obtain for/, instead of the equation A/ = constant,
A/+ k 2 f — constant.
Omitting an unimportant constant term in the solution of this equation, we find/ = ae ikr lr
taking the solution which vanishes at infinity. The constant a is determined from the boundary
condition that v = SI X r at the surface of the sphere. The result is
i? 3 / R\ 3 likr ,, x
/ = e ik(rR) v = (SI xr) — z — ^ Wr_iJ) ,
where R is the radius of the sphere. A calculation like that in §20, Problem 1, gives the fol
lowing expression for the moment of the forces exerted on the sphere by the fluid:
&r 3 + 6R/8+ 6(K/8)2+ 2(i?/3)s 2«(R/S)2(1 +JJ/8)
M ~ ~ T^ l + 2*/8 + 2(W '
For co > (i.e. S > oo), we obtain M = — 8itt)R 3 Q, corresponding to uniform rotation
of the sphere (see §20, Problem 1). In the opposite limiting case RI8 > 1, we find
4a/2
M = ^ir2?V0v>G>)(*  1) Q 
98 Viscous Fluids §25
This expression can also be obtained directly: for S <^ R each element of the surface of the
sphere may be regarded as plane, and the frictional force acting on it is found by substituting
a = OR sin in formula (24.6).
Problem 10. Determine the moment of the forces on a hollow sphere filled with viscous
fluid and executing rotary oscillations about a diameter.
Solution. We seek the velocity in the same form as in Problem 9. For/ we take the solu
tion (a/r) sin kr, which is finite everywhere within the sphere, including the centre. Deter
mining a from the boundary condition, we have
R \ 3 krcoskr—sinkr
v = (Hxr)l
= (ftxr)(
/ kR cos kR — sin kR '
A calculation of the moment of the frictional forces gives the expression
k 2 R* sin kR + 3kR cos kR  3 sin kR
M = l7rr)R 3 Q
kRcoskR—sinkR
The limiting value for S > 1 is of course the same as in the preceding problem. If
R/S <^1 we have
R2co
I R?co \
The first term corresponds to the inertial forces occurring in the rigid rotation of the whole
fluid.
§25. Damping of gravity waves
Arguments similar to those given above can be advanced concerning the
velocity distribution near the free surface of a fluid. Let us consider oscil
latory motion occurring near the surface (for example, gravity waves).
We suppose that the conditions (24.11) hold, the dimension / being now re
placed by the wavelength A:
A2w > v , a<£\; (25.1)
a is the amplitude of the wave, and w its frequency. Then we can say that
the flow is rotational only in a thin surface layer, while throughout the rest
of the fluid we have potential flow, just as we should for an ideal fluid.
The motion of a viscous fluid must satisfy the boundary conditions (15.14)
at the free surface; these require that certain combinations of the space
derivatives of the velocity should vanish. The flow obtained by solving the
equations of idealfluid dynamics does not satisfy these conditions, however.
As in the discussion of % in the previous section, we may conclude that the
corresponding velocity derivatives decrease rapidly in a thin surface layer.
It is important to notice that this does not imply a large velocity gradient as
it does near a solid surface.
Let us calculate the energy dissipation in a gravity wave. Here we must
consider the dissipation, not of the kinetic energy alone, but of the mechanical
energy E mech , which includes both the kinetic energy and the potential
§25 Damping of gravity waves 99
energy in the gravitational field. It is clear, however, that the presence or
absence of a gravitational field cannot affect the .energy dissipation due to
processes of internal friction in the fluid. Hence E mech is given by the same
formula (16.3):
J \ OXjc OXi l
In calculating this integral for a gravity wave, it is to be noticed that, since
the volume of the surface region of rotational flow is small, while the velocity
gradient there is not large, the existence of this region may be ignored, unlike
what was possible for oscillations of a solid surface. In other words, the inte
gration is to be taken over the whole volume of fluid, which, as we have seen,
moves as if it were an ideal fluid.
The flow in a gravity wave for an ideal fluid, however, has already been
determined in §12. Since we have potential flow,
dvijdxk = d 2 <f>/dxjcdxi = Bvjt/dxu
so that
A— a* J '(^k) dV 
The potential <f> is of the form
<j> = <f>ocos(kx—o)t + (x.)e~ kz .
We are interested, of course, not in the instantaneous value of the energy
dissipation, but in its mean value £ mech with respect to time. Noticing that
the mean values of the squared sine and cosine are the same, we find
•Cmech
= 8^4 (<jfidV. (25.2)
The energy E mech itself may be calculated for a gravity wave by using a
theorem of mechanics that, in any system executing small oscillations (of
small amplitude, that is), the mean kinetic and potential energies are equal.
We can therefore write E meeh simply as twice the kinetic energy:
£ me ch = P jv*dV = pj (d<f>ldxi)2 dV,
whence
Em** = 2p& JP dV. (25.3)
The damping of the waves is conveniently characterised by the damping
coefficient y, defined as
y =  E m ech \l2E meC h* (25.4)
100 Viscous Fluids §25
In the course of time, the energy of the wave decreases according to the law
^mech = constant x e~ 2n ; since the energy is proportional to the square of
the amplitude, the latter decreases with time as e~ n .
Using (25.2), (25.3), we find
y = 2vk 2 . (25.5)
Substituting here (12.7), we obtain the damping coefficient for gravity waves
in the form
y = 2vcD*lg*. (25.6)
PROBLEMS
Problem 1. Determine the damping coefficient for long gravity waves propagated in a
channel of constant crosssection; the frequency is supposed so large that ■y/ivloS) is small
compared with the depth of the fluid in the channel.
Solution. The principal dissipation of energy occurs in the surface layer of fluid, where
the velocity changes from zero at the boundary to the value v = v e~ im which it has in the
wave. The mean energy dissipation per unit length of the channel is by (24.14) Hv^^/i^pcolS),
where / is the perimeter of the part of the channel crossjsection occupied by the fluid. The
mean energy of the fluid (again per unit length) is Spv 2 — iSp\vo\ a , where S is the cross
sectional area of the fluid in the channel. The damping coefficient is y = l\Z(vcx)/SS 2 ).
For a channel of rectangular section, therefore,
2h + a
where a is the width and h the depth of the fluid.
Problem 2. Determine the flow in a gravity wave on a very viscous fluid.
Solution. The calculation of the damping coefficient as shown above is valid only when
this coefficient is small, so that the motion may be regarded as that of an ideal fluid to a first
approximation. For arbitrary viscosity we seek a solution of the equations of motion
/ d*V x d*V x \ 1 dp
\ dx 2 dz 2 J p dx'
dv z _ I 8 2 v z 8 2 v z \ I dp
~dt ~ V [~8x^' + ~8z 2 ') "p^  ^
dv x dv z
—  + — =
dx dz
(1)
which depends on t and * as e i(0t + ikx , and diminishes in the interior of the fluid (z < 0).
We find
ik
v x = e  io)t+ikx (Ae kz +Be! mz ), v z = e~ M+ikx (iAe kz Be mz ),
m
p/p = e io>t+ikx coAe kz jk—gz, where m = ^(hP — ico/v).
The boundary conditions at the fluid surface are
(dv x dv z \
1 = for z = I.
dz dx J
§25 Damping of gravity waves 101
In the second condition we can immediately put z = instead of z = £. The first condition,
however, should be differentiated with respect to t, after which we replace gdlldt by gv z
and then put z = 0. The condition that the resulting two homogeneous equations for A
and B are compatible gives
This equation gives w as a function of the wave number k; <o is complex, its real part giving
the frequency of the oscillations and its imaginary part the damping coefficient. The solu
tions of equation (2) that have a physical meaning are those whose imaginary parts are nega
tive (corresponding to damping of the wave); only two roots of (2), meet this requirement.
If vk 2 < V(g k ) ( the condition (25.1)), then the damping coefficient is small, and (2) gives
approximately a> = ± V(gk)i.2vk 2 , a result which we already know. In the opposite limit
ing case vk 2 > V(ik), equation (2) has two purely imaginary roots, corresponding to damped
aperiodic flow. One root is to = igl2vk, while the other is much larger (of order vk 2 ),
and therefore of no interest, since the corresponding motion is strongly damped.
CHAPTER III
TURBULENCE
§26. Stability of steady flow
In solving the equations of steady flow for a viscous fluid, it is often necessary
to make certain approximations on account of mathematical difficulties.
The validity of these approximate solutions is, of course, restricted, Such,
for instance, is the solution of the problem of flow past a sphere given in
§20, which is valid only for small Reynolds numbers.
In principle, however, there must be an exact stationary solution of the
equations of fluid dynamics for any problem with given steady external
conditions; such exact solutions have been considered in §§17, 18 and 23.
These solutions formally hold for all Reynolds numbers.
Yet not every solution of the equations of motion, even if it is exact,
can actually occur in Nature. The flows that occur in Nature must not only
obey the equations of fluid dynamics, but also be stable. For the flow to be
stable it is necessary that small perturbations, if they arise, should decrease
with time. If, on the contrary, the small perturbations which inevitably occur
in the flow tend to increase with time, then the flow is absolutely unstable.
Such a flow unstable with respect to infinitely small perturbations cannot
exist.
The mathematical investigation of the stability of a given flow with respect
to infinitely small perturbations will proceed as follows. On the steady
solution concerned (whose velocity distribution is v (x,y y z), say), we
superpose a nonsteady small perturbation vi (x, y, z, t), which must be
such that the resulting velocity v = v + vi satisfies the equations of motion.
The equation for vi is obtained by substituting in the equations
Sv firad p
— + (vgrad)v = + „Av, divv =
ot p
the velocity and pressure v = v + vi, p = p +p h where the known functions
vo and po satisfy the unperturbed equations
grad/>o
(vograd)vo = + i>Av , divv = 0.
P
Omitting terms above the first order in vi, we obtain
dvi
+ (v • grad)vi + (vi • grad)vo
8t
gradpi
+ vAvi, divvi = 0. (26.1)
P
The boundary condition is that vi vanishes on fixed solid surfaces.
102
§27 The onset of turbulence 103
Thus vi satisfies a system of linear differential equations, with coefficients
that are functions of the coordinates only, and not of the time. The general
solution of such equations can be represented as a sum of particular solutions
in which vi depends on time as e~ ia,t . The "frequencies" to of the perturba
tions are not arbitrary, but are determined by solving the equations (26.1)
with the appropriate boundary conditions. The "frequencies" are in general
complex. If there are w whose imaginary parts are positive, e~ i0it will
increase indefinitely with time. In other words, such perturbations, once
having arisen, will increase, i.e. the flow is unstable with respect to such
perturbations. For the flow to be stable it is necessary that the imaginary
part of any possible "frequency" a> is negative. The perturbations that arise
will then decrease exponentially with time.
Such a mathematical investigation of stability is extremely complicated,
however. The theoretical problem of the stability of steady flow past bodies
of finite dimensions has not yet been solved. It is certain that steady flow is
stable for sufficiently small Reynolds numbers. The experimental data
seem to indicate that, when R increases, it eventually reaches a value Rcr
(the critical Reynolds number) beyond which the flow is unstable with respect
to infinitesimal disturbances. For sufficiently large Reynolds numbers
(R > Rcr), steady flow past solid bodies is therefore impossible. The
critical Reynolds number is not, of course, a universal constant, but takes a
different value for each type of flow. These values appear to be of the order
of 10 to 100; for example, in flow across a cylinder undamped nonsteady
flow has been observed for R = udjv = 34, d being the diameter of the
cylinder. Exact measurements of R C r, however, have not been made.
§27. The onset of turbulence
Let us now consider the nature of the nonsteady flow which is established
as a result of the absolute instability of steady flow at large Reynolds numbers.
We begin by examining the properties of this flow at Reynolds numbers only
slightly greater than R cr . For R < R cr the imaginary parts of the complex
"frequencies" <o = a>i + iyi for all possible small velocity perturbations are
negative (yi < 0). For R = R cr there is one frequency whose imaginary part
is zero. For R > R er the imaginary part of this frequency is positive, but,
when R is close to R cr , yi is small in comparison with the real part wi.f
The function vi corresponding to this frequency is of the form
Vl = A(t)f(x,y,z), (27.1)
where f is some complex function of the coordinates, and the complex
"amplitude" A(t) is$
A(t) = constant KePter*^*. (27.2)
t It must be borne in mind that the set (or spectrum) of all possible frequencies for a given type
of flow includes both separate isolated values (the discrete spectrum) and the whole of various fre
quency ranges (the continuous spectrum). However, it can be seen that the frequencies with positive
imaginary parts in which we are interested occur, in general, Only in the discrete spectrum.
J As usual, we understand the real part of (27.2).
104 Turbulence §27
This expression for A(t) is actually valid, however, only during a short
interval of time after the disruption of the steady flow; the factor e y ^ increases
rapidly with time, whereas the method of determining vi given in §26,
which leads to expressions like (27.1) and (27.2), applies only when jvij
is small. In reality, of course, the modulus \A  of the amplitude of the non
steady flow does not increase without limit, but tends to a finite value.
For R close to R C r (we always mean, of course, R > R cr ), this finite value is
small, and can be determined as follows.
Let us find the time derivative of the squared amplitude \A  2 . For very
small values of t, when (27.2) is still valid, we have d^ 2 /di = 2yi^ 2 .
This expression is really just the first term in an expansion in series of powers
of A and A*. As the modulus \A  increases (still remaining small), sub
sequent terms in this expansion must be taken into account. The next
terms are those of the third order in A. However, we are not interested in
the exact value of the derivative d.4 2 /d/, but in its time average, taken
over times large compared with the period 2ttJo}\ of the factor g^i*; we
recall that, since coi > yi, this period is small compared with the time 1/yi
required for the amplitude modulus \A  to change appreciably. The third
order terms, however, must contain the periodic factor, and therefore vanish
on averaging.f The fourthorder terms include one which is proportional
to A 2 A* 2 = \A  4 and which clearly does not vanish on averaging. Thus we
have as far as fourthorder terms
d\Ap/dt = 2 n \A\*K\A\*. (27.3)
where a may be either positive or negative.
Let us suppose that a is positive. J We have not put bars above \A\ 2
and \A  4 , since the averaging is only over time intervals short compared with
1/yi. For the same reason, in solving the equation we proceed as if the bar
were omitted above the derivative also. The solution of equation (27.3) is
1/^4 2 = a/2yi + constant xr 2 M
Hence it is clear that \A  2 tends asymptotically to a finite limit:
^ 2 max = 2yi/a. (27.4)
The quantity y\ is some function of the Reynolds number. Near R cr it
can be expanded as a series of powers of R— R cr . But yi(R C r) = 0, by
the definition of the critical Reynolds number. Hence the zeroorder term
in the expansion is zero, and we have to the first order y\ = constant x
(R— R C r). Substituting this in (27.4), we see that the modulus \A\ of the
amplitude is proportional to the square root of R— R cr :
^max~ V(RRcr). (27.5)
f Strictly speaking, the thirdorder terms give, on averaging, not zero, but fourthorder terms,
which we suppose included among the fourthorder terms in the expansion.
J This seems to be true for ordinary flow past bodies.
§27 The onset of turbulence 105
Let us summarise these results. The absolute instability of the flow for
R > R cr leads to the appearance of a nonsteady periodic flow. For R close
to Rcr the latter flow can be represented by superposing on the steady flow
vo (x, y, z) a periodic flow vi{x, y, z, t) y with a small but finite amplitude
which increases with R proportionally to the square root of RR er  The
velocity distribution in this flow is of the form
vi = f(*,v,*>*KW, (27.6)
where f is a complex function of the coordinates, and pi is some initial
phase. For large RR cr , the separation of the velocity into v and vi is
no longer meaningful. We then have simply some periodic flow with fre
quency oji. If, instead of the time, we use as an independent variable the
phase <£i s uit+pi, then we can say that the function v(#, y, *, <£i) is a
periodic function of <£i, with period 2tt. This function, however, is no
longer a simple trigonometrical function. Its expansion in Fourier series
V
(where the summation is over all integers p, positive and negative) includes
not only terms with the fundamental frequency a>i, but also terms whose
frequencies are integral multiples of coi.
The following important property of this nonsteady flow should also be
mentioned. Equation (27.3) determines only the modulus of the time factor
A(t), and not its phase. The phase <fc = ant + fii of the periodic flow remains
essentially indeterminate, and depends on the particular initial conditions
which happen to occur at the instant when the flow begins. The initial
phase ft can have any value, depending on these conditions. Thus the
periodic flow under consideration is not uniquely determined by the given
steady external conditions in which the flow takes place. One quantity— the
initial phase of the velocity— remains arbitrary. We may say that the flow
has one degree of freedom, whereas steady flow, which is entirely determined
by the external conditions, has no degrees of freedom.
Let us now consider the phenomena which occur when the Reynolds
number increases further. When this happens, a time finally comes when the
periodic flow discussed above in turn becomes unstable. The investigation of
this instability would proceedf similarly to the method given above for
determining the instability of the original steady flow. The part of the un
perturbed flow is now taken by the periodic flow v (x, y, z, t) (with frequency
wi), and in the equations of motion we substitute v = v + v 2 , where v 2
is a small correction. For v 2 we again obtain a linear equation, but the co
efficients are now functions of time as well as of the coordinates, being
t But has not been carried out even for particular cases, on account of the exceptional mathematical
difficulties.
*06 Turbulence
§27
periodic in time with period 2tt/coi. The solution of such an equation must
be sought in the form v 2 = U(x, y, z, t)er^\ where II(tf, y, z f t) is a periodic
function of time, with period 2tt/o> 1 . The instability again occurs when a fre
quency cu = io 2 + iy 2 appears such that the imaginary part y 2 is positive, and
the corresponding real part o> 2 then determines the new frequency which
appears.
The result, therefore, is that a quasiperiodic flow appears, characterised
by two different periods. Just as the flow had one degree of freedom after
the appearance of the first periodic flow, so it now involves two arbitrary
quantities (phases), i.e. it has two degrees of freedom.
When the Reynolds number increases still further, more and more new
periods appear in succession. The range of Reynolds numbers between
successive appearances of new frequencies diminishes rapidly in size. The
new flows themselves are on a smaller and smaller scale. This means that
the order of magnitude of the distances over which the velocity changes
appreciably is the smaller, the later the flow in question appears.
For R > R cr , therefore, the flow rapidly becomes complicated and con
fused. Such a flow is said to be turbulent; its properties will be investigated
in detail in the following sections. In contradistinction to turbulent flow,
the regular flow, in which the fluid moves as it were in layers with different
velocities, is said to be laminar.
We can write down the general form of a function v(x, y, z, t) whose time
dependence is given by some number n of different frequencies w } (j = 1,
2, ..., «). Instead of one phase fa = a>i*+ft, we now have n different phases
fa = wjt+Pj. The function v may be regarded as a function of these phases
(and of the coordinates), and is periodic in each of them, with period 2tt.
Such a function can be written as a series:
v(x,y,z,t)= £ ^....Pni^yy^explif^p^l (27.8)
Pi'PvPn i=l
the summation being taken over all integrals p h p 2 , ..., p n . This is a generali
sation of formula (27.7). We may notice that the choice of the fundamental
frequencies o>i, ..., co n is, as we see from (27.8), itself not unique; we could
equally well take any n independent linear combinations of co t with integral
coefficients.f
A flow described by a formula such as (27.8) has n degrees of freedom;
it involves n arbitrary initial phases /fy. As the Reynolds number increases,
both the number of frequencies and the number of degrees of freedom
increase. In the limit as R tends to infinity, the number of degrees of free
dom also increases indefinitely.
t These linear combinations must be such that from them we can form all possible numbers
S Pity. It is easy to see that, for this to be so, the determinant of the transformation coefficients
relating the old and new frequencies must be unity.
§28 Stability of flow between rotating cylinders 107
It must be borne in mind that, since the velocity is a periodic function of
the phases, with period 2tt, the states whose phases differ only by an integral
multiple of 2rr are physically indistinguishable. In other words, we can
say that all the essentially different values of each phase lie in the range
s% fa ^ 277. Let us consider any two phases ^i = <oit+pia.ndfa = o>2*+/?2.
Suppose that, at some instant, fa has the value a. Then, by what we have
just said, fa will have values equivalent to a at all instants t = (a  ft. + 27rr)/a>i,
where r is any integer. At these instants the phase fa will have the values
fa = a>2(ajSi)/ft>i+j82 + 277TG)2/c<;i.
The different frequencies are generally incommensurable, so that co 2 /o>i
is an irrational number. If we reduce each value of fa to the range to 2tt
by subtracting the appropriate integral multiple of 2tt, we therefore obtain,
as r goes from to oo, values for fa which are arbitrarily close to any given
number in that range. In other words, in the course of a sufficiently long time
fa and fa will simultaneously be arbitrarily close to any given pair of values.
The same is obviously true of all the phases. Thus turbulent motion has
a certain quasiperiodic property: in the course of a sufficiently long time the
fluid passes through states arbitrarily close to any given state, determined by
any possible choice of simultaneous values of the phases fa.
We have introduced the concept of the critical Reynolds number as being
the value of R at which instability of steady flow, in the sense described above,
first occurs. The critical Reynolds number can, however, be regarded from
a somewhat different point of view. For R < R cr there are no stable non
steady solutions of the equations of motion that are not damped in time.
After the critical value has been reached, a stable nonsteady solution appears,
which will actually occur in a moving fluid.
As far as experimental investigations of the flow past ordinary finite
bodies are concerned, the two definitions of R cr seem to be the same. Logi
cally, however, this need not be so, and cases could in principle occur where
there are two different critical values: one above which nonsteady flow
can occur without being damped, and another above which steady flow
becomes absolutely unstable. The second must obviously be greater than
the first. However, since there is at present no indication that such cases of
instability actually exist, we shall not pause to investigate them more closely.f
§28. Stability of flow between rotating cylinders
To investigate the stability of steady flow between two rotating cylinders
(§18) in the limit of very large Reynolds numbers, we can use a simple method
like that used in §4 to derive the condition for mechanical stability of a fluid
at rest in a gravitational field (Rayleigh, 1916). The principle of the method
is to consider any small element of the fluid and to suppose that this element
t We are not here concerned with (e.g.) flow in a pipe, where the loss of stability has unusual
properties (see §29).
108 Turbulence §28
is displaced from the path which it follows in the flow concerned. As a result
of this displacement, forces appear which act on the displaced element. If
the original flow is stable, these forces must tend to return the element to
its original position.
Each fluid element in the unperturbed flow moves in a circle r = constant
about the axis of the cylinders. Let fju(r) = tnr 2 <j> be the angular momentum
of an element of mass m, <j> being the angular velocity. The centrifugal
force acting on it is y?\mr z \ this force is balanced by the radial pressure
gradient in the rotating fluid. Let us now suppose that a fluid element at a
distance r from the axis is slightly displaced from its path, being moved to
a distance r > r from the axis. The angular momentum of the element
remains equal to its original value fj. = /i(r ). The centrifugal force acting
on the element in its new position is therefore /xo 2 /wr 3 . In order that the
element should tend to return to its initial position, this force must be less
than the equilibrium value /x 2 /wr 3 which is balanced by the pressure gradient
at the distance r. Thus the necessary condition for stability is /x 2 /x 2 > 0.
Expanding /x(r) in powers of the positive difference rr , we can write this
condition in the form
ndfi/dr > 0. (28.1)
According to formula (18.3), the angular velocity <j> of the moving fluid
particles is
_ Q 2 R2 2 niRi 2 (Qin 2 )Ri 2 R2 2 1
R 2 2 ~Ri 2 + R 2 2 Ri 2 r2*
Calculating //, = mr 2 <f> and omitting factors which are certainly positive,
we can write the condition (28.1) as
(Q 2 #2 2 ai#i 2 >£ > 0. (28.2)
The angular velocity <f> varies monotonically from Qi on the inner cylinder
to Q 2 on the outer cylinder. If the two cylinders rotate in opposite directions,
i.e. if £li and Q2 have opposite signs, the function (/> changes sign between the
cylinders, and its product with the constant number D 2 i2 2 2 ^ii?i 2 cannot
be everywhere positive. Thus in this case (28.2) does not hold at all points
in the fluid, and the flow is unstable.
Now let the two cylinders be rotating in the same direction; taking this
direction of rotation as positive, we have Qi > 0, Q 2 > 0. Then </> is every
where positive, and for the condition (28.2) to be fulfilled it is necessary that
Q 2 i?2 2 > Q1.R1 2 . (28.3)
If Q2R2 2 < Q.1R1 2 the flow is unstable. For example, if the outer cylinder is
at rest (Q 2 = 0), while the inner one rotates, then the flow is unstable. If,
on the other hand, the inner cylinder is at rest (Qi = 0), the flow is stable.
It must be emphasised that no account has been taken, in the above argu
ments, of the effect of the viscous forces when the fluid element is displaced.
§28 Stability of flow between rotating cylinders 109
The method is therefore applicable only for small viscosities, i.e. for large R.
To investigate the stability of the flow for any R, it is necessary to follow
the general method, starting from equations (26.1) (G. I. Taylor, 1923).
In the present case the unperturbed velocity distribution vo depends only on
the (cylindrical) radial coordinate r, and not on the angle <f> or the axial
coordinate z. Thus we have for the perturbation vi a system of linear
equations with coefficients which contain neither the time nor the coordinates
<f> and z. We may seek solutions of these equations in the form
Vl _ e iVcz<ot)f( r ) f (28.4)
the direction of the vector f being arbitrary; this solution depends on z
through the periodic factor e ikz , and the wave number k determines the
periodicity of the perturbation in the zdirection. The possible frequencies
co, obtained by solving the equations with the necessary boundary conditions
in a plane perpendicular to the axis (vi = for r = Ri and r = R 2 ), will
then be functions of k, involving R as a parameter: co = co(k, R). The
point where instability appears is determined by the value of R for which the
function y\ = im co first becomes zero for some k. For R < R C r, the func
tion yi(k, R) is always negative, but for R > R cr we have y 1 > in some range
of k. Let k cr be the value of k for which yi = when R = R cr  The cor
responding function (28.4) gives the nature of the flow which occurs (super
posed on the original flow) in the fluid at the instant when the original flow
ceases to be stable; it is periodic along the axis of the cylinders, with wave
length 277/& cr .t
As well as solutions of the form (28.4), which are independent of the angle
<f>, the system of equations under consideration has also solutions for which
vi contains a factor e' m *, m being an integer. We are, however, interested
only in the solution which corresponds to the first appearance of instability.
The solutions with m # have never been studied in this respect. It is
nevertheless natural to suppose that instability occurs first of all with respect
to perturbations with m = 0, a supposition which is entirely confirmed by
experimental results.
It should also be borne in mind that, even for a given k, the solution of the
form (28.4) is not unique. In general, a number of solutions with different
values of co correspond to a given k. Again we are interested only in the one
which gives the smallest value of R C r
It is found that a purely imaginary function co(k) corresponds to the solu
tion which gives the smallest R cr . Hence, when k = k cx , not only im co
but co itself is zero. This means that the first instability of the steady flow
between rotating cylinders leads to the appearance of another flow which is
also steady.
t For R slightly greater than R cr there is not one value of k, but a whole range, for which im a> > 0.
However, it should not be thought that the resulting flow will be a superposition of flows with various
periodicities. In reality, for each R a flow of definite periodicity occurs which stabilises the total
flow. This periodicity, however, cannot be determined from the linearised equations (26.1).
110
Turbulence
§28
On account of the great complexity of the calculation,! numerical results
have been obtained only for the case where the space between the cylinders
is narrow {R 2 R\ <^ R2). Fig. 13 shows an example of the curve separating
the regions of unstable (shaded) and stable flow. The righthand branch of
the curve, corresponding to rotation of the two cylinders in the same direc
tion, is asymptotic to the line Q 2 #2 2 = Hii?i 2 . When the Reynolds number
increases, for a given type of flow, the two numbers Cltf/jv and Q 2 #2 2 /v
increase by equal factors. In Fig. 13 this corresponds to a movement upwards
along a line through the origin having a given slope. In the righthand part
of the diagram (Di and Q 2 both positive), such lines for which a 2 R2 2 l&iRi 2 > 1
do not meet the curve which bounds the region of instability. If, on the other
hand, Q 2 .R 2 2 /fti.Ri 2 < 1, then for sufficiently large Reynolds numbers we
enter the region of instability, in accordance with the condition (28.3).
Fig. 13
In the lefthand part of the diagram (Qi and Q 2 of opposite signs), any line
through the origin eventually meets the curve, i.e. the flow can become un
stable for any value of the ratio £l 2 R2 2 l&iRi 2 , again in agreement with the
results obtained above. For Q 2 = (when only the inner cylinder rotates),
instability sets in when
Oi = A\Zvlhy/{hR 2 \ (28.5)
where h = R 2 Ri.
The stability of the flow in the unshaded part of Fig. 13 does not mean,
however, that the flow actually remains steady no matter how large R be
comes. Experiment shows that there is a limit beyond which stable non
steady flow becomes possible. In this region the steady flow is "metastable" :
it is stable with respect to small perturbations, but unstable with respect to
larger perturbations. If, owing to such perturbations, nonsteady flow occurs
in some region along the cylinders, it will subsequently "displace" the laminar
flow in all space. This nonsteady flow has, as soon as it appears, a large
number of "degrees of freedom" (in the sense explained in §27), i.e. it is
fully developed turbulence.
t Further details may be found in the book by C. C. LtN, The Theory of Hydrodynamic Stability,
Cambridge 1955.
§29 Stability of flow in a pipe 111
In the shaded part of Fig. 13, the flow again becomes turbulent for
sufficiently large R, but there are, it seems, very few data concerning the
way in which it appears.
A limiting case of the flow between rotating cylinders, corresponding to
large radii and small h = R2R1, is flow between two parallel planes in
relative motion (see §17). This flow is stable with respect to infinitely
small perturbations for any value of R = Uhjv, where U is the relative
velocity of the planes. Stable turbulent motion becomes possible, however,
for values of R greater than about 1500.
§29. Stability of flow in a pipe
The steady flow in a pipe discussed in §17 loses its stability in an unusual
manner. Since the flow is uniform in the ^direction (along the pipe), the
unperturbed velocity distribution vo is independent of x. Similarly to the
procedure in §28, we can therefore seek solutions of equations (26.1) in the
form
Vl = ««*»*> f[y,*). (29.1)
Here also there is a value R = R cr for which yi = im w first becomes zero
for some value of k. It is of importance, however, that the real part of the
function co(k) is not now zero.
R>R £
R=R C
R<R«
Fig. 14
For values of R only slightly exceeding R cr , the range of values of k for
which yi(k) > is small and lies near the point for which yi(k) is a maximum,
i.e. dyi/dk = (as seen from Fig. 14). Let a slight perturbation occur in
some part of the flow; it is a wave packet obtained by superposing a series of
components of the form (29.1). In the course of time, the components for
which y\(k) > will be amplified, while the remainder will be damped.
The amplified wave packet thus formed will also be carried downstream with
a velocity equal to the group velocity dcojdk of the packet; since we are now
considering waves whose wave numbers lie in a small range near the point
where dyijdk = 0, the quantity dco/dk « dcoi[dk is real, and is therefore the
actual velocity of propagation of the packet.
112 Turbulence §29
This downstream displacement of the perturbations is very important, and
causes the loss of stability to be totally different from that described in §28.
We have seen that, for flow between rotating cylinders with R > R cr
(when there are frequencies with im o> > 0), the original steady flow is no
longer possible, since even small perturbations are increased to a finite
amplitude. For flow in a pipe, however, the amplification of the perturbation
is accompanied by its displacement downstream ; if we consider the flow at a
given point in the pipe, it is found that the perturbation there is not amplified,
but damped. It must also be borne in mind that, since in reality we have
pipes of finite length, however great, any perturbation may be carried out of
the pipe before it disrupts the laminar flow. Thus, even for R > R cr , steady
flow in a pipe is effectively stable with respect to small perturbations, and can
in principle take place for values of R considerably exceeding R cr .
Since the perturbations increase with the coordinate x (downstream),
and not with time at a given point, it is reasonable to investigate this type of
instability as follows. Let us suppose that, at a given point, a continuously
acting perturbation with a given frequency co is applied to the flow, and
examine what will happen to this perturbation as it is carried downstream.
Inverting the function o> = a)(k), we find what wave number k corresponds
to the given (real) frequency co. If im k < 0, the factor e ikx increases with
x, i.e. the perturbation is amplified downstream. The curve in the tuRplane
given by the equation im k(o), R) = defines the region of stability, and
separates, for each R, the frequencies of perturbations which are amplified
and damped downstream.
The actual calculations are extremely complicated. A complete investi
gation has been made only for flow between two parallel planes (C. C. Lin,
1946).f However, it is reasonable to suppose that the results will be quali
tatively the same for flow in a circular pipe.
The limiting curve for flow between two planes is schematically shown in
Fig. 15. The shaded area within the curve is the region of instability. As
R > oo, both branches of the curve are asymptotic to the Raxis.J For
the smallest value of R at which undamped perturbations are possible we
find by calculations R cr « 7700, R being defined as Uhjv, with h the distance
between the planes and U the fluid velocity averaged across this distance.
Thus, for any frequency between zero and a certain maximum value, there
is a finite range of R values for which perturbations with the frequency con
cerned will be amplified. It is interesting to note that a small but finite
viscosity of the fluid has, in a sense, a destabilising effect in comparison with
the situation for a strictly ideal fluid. For, when R > oo, perturbations with
any finite frequency are damped, but when a finite viscosity is introduced we
eventually reach a region of instability; a further increase in the viscosity
(decrease in R) finally brings us out of this region.
f A detailed account is given by C. C. Lin, The Theory of Hydrodynamic Stability, Cambridge 1955.
{ The asymptotic equations of the two branches for large R are a>A/C7 = 50/R a/l1 , coh/U = 112/R 3 / 7 .
§29
Stability of flow in a pipe
113
These calculations, however, do not answer the question whether, for
sufficiently large R, flow in a pipe does not also exhibit true instability with
respect to infinitely small perturbations, i.e. instability resulting in the
amplification of perturbations with time at a given point. We shall outline
the mathematical significance of such an instability. Let us consider some
small perturbation which occurs at time t = in a finite region. Expanding
it as a Fourier integral with respect to x, we can write it as
l\M)e ik{x ^dk>
where f(x) is a function describing the initial perturbation. In the course of
time, each Fourier component of the perturbation will vary as e ia>t , with a
frequency u> = a>{k, R), so that the whole perturbation at time t will be given
by the integral
\(f(tj)e mx ® i0)t d{;dk.
Since f(x) is zero except in a finite region, x  has a finite range of values.
Hence the behaviour of the integral for large t is essentially determined by
the behaviour of the integral
f e iaUc)tdk.
If this integral tends to infinity with t f the flow is in fact absolutely unstable.
Fig. IS
No such investigation has yet been made, even for a particular case.
However, the experimental results concerning flow in pipes give reason to
suppose that there is no true instability with respect to arbitrarily small
perturbations for any R. This is indicated by the fact that, the more care
fully perturbations at the entrance to the pipe are prevented, the larger the
Reynolds numbers for which laminar flow can be observed.f
t Laminar flow has actually been observed up to R« 50,000, where R = Ud[v, d being the diameter
of the pipe and U the mean velocity over its crosssection.
114 Turbulence §30
However, the experimental results also show that there is another critical
Reynolds number (which we denote by R c /); this determines the limit
beyond which stable nonsteady flow can exist (cf. the end of §27). If, in
any section of the pipe, turbulent flow occurs, then for R < R cr ' the turbulent
region will be carried downstream and will diminish in size until it disappears
altogether; if, on the other hand, R > R cr ', the turbulent region will extend
in the course of time to include more and more of the flow. If perturbations
of the flow occur continually at the entrance to the pipe, then for R < R cr '
they will be damped out at some distance down the pipe, no matter how strong
they are initially. If, on the other hand, R > R cr ', the flow becomes turbulent
throughout the pipe, and this can be achieved by perturbations which are the
weaker, the greater R.f Thus laminar flow in a pipe with R > R cr ' is
metastable, being unstable with respect to perturbations of finite intensity;
the necessary intensity is the smaller, the greater R.
As has been mentioned at the end of §28, nonsteady flow arising by the
disruption of metastable laminar flow is already fullydeveloped turbulence.
In this sense the appearance of turbulence in a pipe is essentially different
from the appearance of turbulence owing to the absolute instability of steady
flow past finite bodies. In the latter case nonsteady flow seems to appear
in a continuous manner as we pass through R cr , the number of degrees of
freedom increasing gradually (as explained in §§26 and 27). For flow in a
pipe, however, turbulence appears discontinuously. This difference causes,
in particular, the different dependence of the drag on the Reynolds number
in the two cases. For example, if we consider the motion of any body in a
fluid, the drag force F on it is continuous at R = R cr , where steady flow
becomes absolutely unstable. At this point the curve F = F(R) can have
only a bend corresponding to the change in the nature of the flow. For
flow in a pipe, on the other hand, there are essentially two different laws of
drag for R > R cr : one for steady flow, and the other for turbulent flow.
The drag is discontinuous for whatever value of R marks the transition from
one type of flow to the other.
§30. Instability of tangential discontinuities
Flows in which two layers of incompressible fluid move relative to each
other, one "sliding" on the other, are absolutely unstable if the fluid is ideal;
the surface of separation between these two fluid layers would be a surface of
tangential discontinuity, on which the fluid velocity tangential to the surface
is discontinuous. We shall see below (§35) what is the actual nature of the
flow resulting from this instability; here we shall prove the above statement.
If we consider a small portion of the surface of discontinuity and the flow
near it, we may regard this portion as plane, and the fluid velocities vi and
V2 on each side of it as constants. Without loss of generality we can suppose
t For a pipe of circular crosssection R cr ' lies between 1600 and 1700. For flow between parallel
planes, turbulent flow has been observed from R = 1400 upwards.
§30 Instability of tangential discontinuities 115
that one of these velocities is zero ; this can always be achieved by a suitable
choice of the coordinate system. Let V2 = 0, and vi be denoted by v
simply ; we take the direction of v as the ffaxis, and the saxis along the normal
to the surface.
Let the surface of discontinuity receive a slight perturbation, in which all
quantities — the coordinates of points on the surface, the pressure, and the
fluid velocity — are periodic functions, proportional to e i{kx ~° >t) . We consider
the fluid on the side where its velocity is v, and denote by v' the small change
in the velocity due to the perturbation. According to the equations (26.1)
(with constant vo = v and v — 0), we have the following system of equations
for the perturbation v' :
dv' grad^)'
diw' = 0, + (v.grad)v' = .
8t p
Since v is along the #axis, the second equation can be rewritten
dv' eV gradp'
f v = . (JU.l)
dt dx p
If we take the divergence of both sides, then the lefthand side gives zero by
virtue of diw' = 0, so that p' must satisfy Laplace's equation:
AP' = 0. (30.2)
Let £ = Z,(x> t) be the displacement in the srdirection of points on the
surface of discontinuity, due to the perturbation. The derivative dt,\dt
is the rate of change of the surface coordinate £ for a given value of x.
Since the fluid velocity component normal to the surface of discontinuity
is equal to the rate of displacement of the surface itself, we have to the
necessary approximation
dl\dt = v'zvdt/dx, (30.3)
where, of course, the value of v' z on the surface must be taken.
We seek p' in the form p' = f(z) *«**«*>. Substituting in (30.2), we
have for f(z) the equation d 2 //ds 2  k 2 f = 0, whence / = constant xe** 2 .
Suppose that the space on the side under consideration (side 1) corresponds to
positive values of z. Then we must take/ = constant xr* z , so that
p' = constant x «*****> <r* z . (30.4)
Substituting this expression in the ^component of equation (30.1), we findf
v'z = kp'xjip^kvoi). (30.5)
The displacement £ may also be sought in a form proportional to the same
exponential factor e t(Jex ~ cot \ and we obtain from (30.3) v' z = i£(kv — a).
t The case kv = co, though possible in principle, is not of interest here, since instability can arise
only from complex frequencies to, not from real a).
116 Turbulence §31
This gives, instead of (30.5),
P'l = t P i{kva>flk. (30.6)
The pressure p\ on the other side of the surface is given by a similar formula,
where now v = and the sign is changed (since in this region z < 0, and all
quantities must be proportional to e kz , not e~ kz ). Thus
p' 2 = faafijk. (30.7)
We have written different densities p\ and p2 in order to include the case
where we have a boundary separating two different immiscible fluids.
Finally, from the condition that the pressures p\ and p\ are equal on
the surface of discontinuity, we obtain pi(kv — co) 2 = — p 2 a) 2 , from which
the desired relation between co and k is found to be
Pi± W(pm) , on ox
o) = kv . (30.8)
P1+P2
We see that o> is complex, and there are always co having a positive imagi
nary part. Thus tangential discontinuities are unstable, even with respect
to infinitely small perturbations. In this form the result is true for very
small viscosities, i.e. for very large R. In this case it is meaningless to
distinguish instability of the type that is "carried along" from true absolute
instability, since, as k increases, the imaginary part of o> increases without
limit, and hence the "amplification coefficient" of the perturbation as it is
carried along may be as large as we please.
When finite viscosity is taken into account, the tangential discontinuity is
no longer sharp; the velocity changes from one value to another across a
layer of finite thickness. The problem of the stability of such a flow is
mathematically entirely similar to that of the stability of flow in a laminar
boundary layer with a point of inflexion in the velocity profile (§41). The
experimental results indicate that instability sets in very soon.
§31. Fully developed turbulence
Turbulent flow at fairly large Reynolds numbers is characterised by the
presence of an extremely irregular variation of the velocity with time at each
point. This is called fully developed turbulence. The velocity continually
fluctuates about some mean value, and it should be noted that the amplitude
of this variation is in general not small in comparison with the magnitude of
the velocity itself. A similar irregular variation of the velocity exists between
points in the flow at a given instant. The paths of the fluid particles in turbu
lent flow are extremely complicated, resulting in an extensive mixing of the
fluid.
As has been mentioned in the previous section, turbulent flow has a very
large number of degrees of freedom. The values of the initial phases fa
corresponding to these degrees of freedom are determined by the initial
§31 Fully developed turbulence 117
conditions of the flow. The specification of the exact initial conditions
which would determine the value of so many quantities is, however, so un
realistic that even to put the problem in this form is physically meaningless.
The position here is similar to what would happen if we attempted to
consider the motion of all the molecules forming a macroscopic body, using
the equations of mechanics; here again the problem of specifying the initial
conditions which determine the initial values of the coordinates and velocities
of all the molecules, and then integrating the equations of motion, is physically
meaningless. The analogy extends further. A macroscopic body, regarded
as composed of individual molecules, has an enormous number of degrees
of freedom. An exact microscopic description of the state of the body would
involve a determination of the coordinates and velocity of every particle
composing it. The exact manner in which these quantities vary with time
depends on their values at the initial instant. However, owing to the extreme
complexity and irregularity of the motion of the molecules, we may suppose
that, over a sufficiently long interval of time, the velocities and coordinates of
the molecules take all possible sets of values, so that the effect of the initial
conditions is smoothed out and disappears. This, as is well known, makes
possible a statistical discussion of macroscopic bodies.
A similar situation occurs in turbulent flow. For an exact description of
the time variation of the velocity distribution in the moving fluid, the values
of all the initial phases /?/ would have to be given; the values of all the phases
<f>j = ajjt + fa at every instant would then be known. We have seen that, what
ever the initial phases /fy, over a sufficiently long interval of time the fluid
passes through states arbitrarily close to any given state, defined by any
possible choice of simultaneous values of the phases <f>j. Hence it follows that,
in the consideration of turbulent flow, the actual initial conditions cease to
have any effect after sufficiently long intervals of time. This shows that the
theory of turbulent flow must be a statistical theory. No complete quantitative
theory of turbulence has yet been evolved. Nevertheless, several very impor
tant qualitative results are known, and the following sections give an account
of these.
We introduce the concept of the mean velocity, obtained by averaging over
long intervals of time the actual velocity at each point. By such an averaging
the irregular variation of the velocity is smoothed out, and the mean velocity
varies smoothly from point to point. In what follows we shall denote the mean
velocity by u = v. The difference v' = v— u between the true velocity
and the mean velocity varies irregularly in the manner characteristic of tur
bulence ; we shall call it the fluctuating part of the velocity.
Let us consider in more detail the nature of this irregular motion which is
superposed on the mean flow. This motion may in turn be qualitatively
regarded as the superposition of turbulent eddies of different sizes ; by the size
of an eddy we mean the order of magnitude of the distances over which the
velocity varies appreciably. As the Reynolds number increases, large eddies
appear first; the smaller the eddies, the later they appear. For very large
118 Turbulence §31
Reynolds numbers, eddies of every size from the largest to the smallest are
present. An important part in any turbulent flow is played by the largest
eddies, whose size is of the order of the dimensions of the region in which the
flow takes place ; in what follows we shall denote by / this order of magnitude
for any given turbulent flow. These large eddies have the largest amplitudes.
The velocity in them is comparable with the variation of the mean velocity
over the distance /; we shall denote by Am the order of magnitude of this
variation.) The frequencies corresponding to these eddies are of the order
of ujl, the ratio of the mean velocity u (and not its variation Am) to the dimen
sion /. For the frequency determines the period with which the flow pattern
is repeated when observed in some fixed frame of reference. Relative to such
a system, however, the whole pattern moves with the fluid at a velocity of the
order of w.
The small eddies, on the other hand, which correspond to large frequencies,
participate in the turbulent flow with much smaller amplitudes. They may
be regarded as a fine detailed structure superposed on the fundamental large
turbulent eddies. Only a comparatively small part of the total kinetic energy
of the fluid resides in the small eddies.
From the picture of turbulent flow given above, we can draw a conclusion
regarding the manner of variation of the fluctuating velocity from point to
point at any given instant. Over large distances (comparable with /), the
variation of the fluctuating velocity is given by the variation in the velocity
of the large eddies, and is therefore comparable with Am. Over small
distances (compared with /), it is determined by the small eddies, and is
therefore small (compared with Am)4 The same kind of picture is obtained if
we observe the variation of the velocity with time at any given point. Over
short time intervals (compared with T ~ l/u), the velocity does not vary
appreciably; over long intervals, it varies by a quantity of the order of Am.
The length / appears as a characteristic dimension in the Reynolds number
R, which determines the properties of a given flow. Besides this Reynolds
number, we can introduce the qualitative concept of the Reynolds numbers
for turbulent eddies of various sizes. If A is the order of magnitude of the
size of a given eddy, and v\ the order of magnitude of its velocity, then
the corresponding Reynolds number is defined as R^ ~ v\^l v  This number
is the smaller, the smaller the size of the eddy.
For large Reynolds numbers R, the Reynolds numbers R\ of the large
eddies are also large. Large Reynolds numbers, however, are equivalent to
small viscosities. We therefore conclude that, for the large eddies which are
the basis of any turbulent flow, the viscosity is unimportant and may be
t We are speaking here of the order of magnitude, not of the mean velocity itself, but of its variation
(over distances of the order of /)> since it is this variation Au which characterises the velocity of the
turbulent flow. The mean velocity itself can have any magnitude, depending on the frame of reference
used.
It may also be mentioned that experimental results indicate that the size of the largest eddies is
actually somewhat less than I, and their velocity is somewhat less than Am.
J But large compared with the variation of the mean velocity over these small distances.
§31 Fully developed turbulence 119
equated to zero, so that the motion of these eddies obeys Euler's equation. In
particular, it follows from this that there is no appreciable dissipation of
energy in the large eddies.
The viscosity of the fluid becomes important only for the smallest eddies,
whose Reynolds number is comparable with unity. We denote the size of
these eddies by A , which we shall determine in the next section. It is in
these small eddies, which are unimportant as regards the general pattern of
a turbulent flow, that the dissipation of energy occurs.
We thus arrive at the following conception of energy dissipation in turbu
lent flow. The energy passes from the large eddies to smaller ones, practi
cally no dissipation occurring in this process. We may say that there is
a continuous flow of energy from large to small eddies, i.e. from small to
large frequencies. This flow of energy is dissipated in the smallest eddies,
where the kinetic energy is transformed into heat.f
Since the viscosity of the fluid is important only for the smallest eddies,
we may say that none of the quantities pertaining to eddies of sizes A > Ao
can depend on v (more exactly, these quantities cannot be changed if v
varies but the other conditions of the motion are unchanged). This circum
stance reduces the number of quantities which determine the properties of
turbulent flow, and the result is that similarity arguments, involving the dimen
sions of the available quantities, become very important in the investigation
of turbulence.
Let us apply these arguments to determine the order of magnitude of the
energy dissipation in turbulent flow. Let e be the mean dissipation of
energy per unit time per unit mass of fluid.} We have seen that this
energy is derived from the large eddies, whence it is gradually transferred to
smaller eddies until it is dissipated in eddies of size ~ Ao. Hence, although
the dissipation is ultimately due to the viscosity, the order of magnitude of e
can be determined only by those quantities which characterise the large
eddies. These are the fluid density p, the dimension / and the velocity
Am. From these three quantities we can form only one having the dimensions
of e, namely erg/g sec = cm 2 /sec 3 . Thus we find
e ~ (Am)3//, (31.1)
and this determines the order of magnitude of the energy dissipation in turbu
lent flow.
In some respects a fluid in turbulent motion may be qualitatively described
as having a "turbulent viscosity" j/ turb which differs from the true kinematic
viscosity v. Since v turb characterises the properties of the turbulent flow,
its order of magnitude must be determined by p, Aw and /. The only quantity
that can be formed from these and has the dimensions of kinematic viscosity
t For a steady state to be maintained, it is of course necessary that external energy sources should
be present which continually supply energy to the large eddies.
X In this chapter e denotes the mean dissipation of energy, and not the internal energy of the
fluid.
120 Turbulence §32
is /Am, and therefore
vturb ~ /Aw. (31.2)
The ratio of the turbulent viscosity to the ordinary viscosity is consequently
v tuTbl v ~ R> * e « i* ^creases with the Reynolds number. f
The energy dissipation e is expressed in terms of v turb by
e  vturb(A M //)2 (31.3)
in accordance with the usual definition of viscosity. Whereas v determines
the energy dissipation in terms of the space derivatives of the true velocity,
v turb re l a tes it to the gradient ( ~ Aw//) of the mean velocity.
We may also apply similarity arguments to determine the order of mag
nitude Ap of the variation of pressure over the region of turbulent flow.
The only quantity having the dimensions of pressure which can be formed
from p, I and Aw is p(Au) 2 . Hence we must have
Ap ~ P (Au)K (31.4)
§32. Local turbulence
Let us now consider the properties of the turbulence as regards eddy sizes
A which are small compared with the fundamental eddy size /. We shall
refer to these properties as local properties of the turbulence. We shall
consider fluid that is far from all solid surfaces (more precisely, that is at
distances from them large compared with A).
It is natural to assume that such smallscale turbulence, far from solid
bodies, is isotropic. This means that, over regions whose dimensions are
small compared with /, the properties of the turbulent flow are independent
of direction ; in particular, they do not depend on the direction of the mean
velocity. It must be emphasised that here, and everywhere in the present
section, when we speak of the properties of the turbulent flow in a small region
of the fluid, we mean the relative motion of the fluid particles in that region,
and not the absolute motion of the region as a whole, which is due to the larger
eddies.
It is found that several very important results concerning the local pro
perties of turbulence can be obtained immediately from similarity arguments.
These results are due to A. N. Kolmogorov and to A. M. Obukhov (1941).
To obtain them, we shall first determine which parameters can be involved in
the properties of turbulent flow over regions small compared with / but large
compared with the distances Xq at which the viscosity of the fluid begins to be
important. It is these intermediate distances which we shall discuss below. The
f In reality, however, a fairly large numerical coefficient should be included. This is because, as
mentioned above, / and Am may differ quite considerably from the actual scale and velocity of the
turbulent flow. The ratio nurb/" may be more accurately written vturb/y ~ R/R<sr> which formula
takes into account the fact that vturb and v must in reality be comparable in magnitude not for R ~ 1,
but for R ~ R cr .
§32 Local turbulence 121
parameters in question are the fluid density p and another quantity charac
terising any turbulent flow, the energy e dissipated per unit time per unit mass
of fluid. We have seen that e is the "energy flux" which continually passes
from larger to smaller eddies. Hence, although the energy dissipation is
ultimately due to the viscosity of the fluid and occurs in the smallest eddies,
the quantity € is determined by the properties of larger eddies. It is natural
to suppose that (for given p and e) the local properties of the turbulence are
independent of the dimension / and velocity Am of the flow as a whole. The
fluid viscosity v also cannot appear in any of the quantities in which we are
at present interested (we recall that we are concerned with distances A > Ao).
Let us determine the order of magnitude v\ of the turbulent velocity varia
tion over distances of the order of A. It must be determined only by p, e
and, of course, the distance A itself. From these three quantities we can
form only one having the dimensions of velocity, namely (eA)*. Hence we
can say that the relation
v x ~ (eA)* (32.1)
must hold. We thus reach a very important result : the velocity variation over
a small distance is proportional to the cube root of the distance {Kolmogorov
and Obukhov's law). The quantity v\ may also be regarded as the velocity
of turbulent eddies whose size is of the order of A.f
Let us now put the problem somewhat differently, and determine the order
of magnitude v 7 of the velocity variation at a given point over a time interval
t which is short compared with the time T ~ lju characterising the flow as
a whole. To do this, we notice that, since there is a net mean flow, any given
portion of the fluid is displaced, during the interval t, over a distance of
the order of ru, u being the mean velocity. Hence the portion of fluid which
is at a given point at time t will have been at a distance ru from that point
at the initial instant. We can therefore obtain the required quantity v r by
direct substitution of ru for A in (32.1):
v T ~ (era)*. (32.2)
Thus the velocity variation over a time interval t is proportional to the cube
root of the interval.
t The variation v x of the velocity over small distances is fundamentally the variation in the fluc
tuating part of the velocity; the variation of the mean velocity over small distances is small compared
with the variation of the fluctuating velocity over those distances.
The relation (32.1) may be obtained in another way by expressing a constant quantity, the dis
sipation e, in terms of quantities characterising the eddies of size A; e must be proportional to the
squared gradient of the velocity v x and to the appropriate turbulent viscosity coefficient
fturb.A ~ C A^
(cf. (31.2), (31.3)):
€ ~ "turb,A(WA) 2 ~ *> A 3 /A,
whence we obtain (32.1).
1 22 Turbulence §32
The quantity v T must be distinguished from v r \ the variation in velocity
of a portion of fluid as it moves about. This variation can evidently depend
only on p and e, which determine the local properties of the turbulence, and
of course on t itself. Forming the only combination of p, e and t that has
the dimensions of velocity, we obtain
v T ' ~ (er)*. (32.3)
Unlike the velocity variation at a given point, it is proportional to the square
root of t, not to the cube root. It is easy to see that, for r small compared
with T, v T ' is always less than v r .\
Using the expression (31.1) for e, we can rewrite (32.1) as
v x ~ Am(A//)*. (32.4)
Similarly, we can write v T as
v T ~ A^r/T)*, (32.5)
where T ~ Iju.
Let us now find at what distances the fluid viscosity begins to be important.
These distances Ao also determine the order of magnitude of the size of the
smallest eddies in the turbulent flow (called the "internal scale" of the tur
bulence, in contradistinction to the "external scale" /). To determine A ,
we form the Reynolds number R A ~ v x \jv\ using (32.4), we obtain
R A ~ AmA4/3/ v /i/3.
Introducing the Reynolds number R ~ /Am/i/ for the flow as a whole, we
can rewrite this as R A ~ R(X/lf. The order of magnitude of A is that for
which R A ~ 1. Hence we find
Ao  //R*. (32.6)
The same expression can be obtained by forming from />, e and v the only
combination having the dimensions of length, namely Ao ~ (i^/e)*, and expres
sing e in terms of Am and / by means of (31.1).
Thus the internal scale of the turbulence is inversely proportional to R f .
For the corresponding velocity we have
v Xa  Am/R*; (32.7)
this also decreases when R increases. Finally, the order of magnitude of the
frequencies corresponding to eddies of this size is too ~ w/Ao or
coo ~ kR*/J. (32.8)
This gives the order of magnitude of the upper end of the frequency spectrum
of the turbulence; the lower end is at frequencies of the order of «//. Thus
the frequency range increases with Reynolds number as R*.
f The inequality v ' <^ v has in essence been assumed in the derivation of (32.2).
§33 The velocity correlation 123
Similar arguments enable us to determine the order of magnitude of the
number of degrees of freedom of a turbulent flow. Let us denote by n the
number of degrees of freedom per unit volume of the fluid ; n has the dimen
sions 1/cm 3 . This number can depend only on p, e and also the viscosity v,
since the latter determines the lower limit of the sizes of the turbulent eddies.
From these three quantities we can form only one having the dimensions
1/cm 3 , namely (e/v 3 )*; this is just 1/Ao 3 , a result which might have been expec
ted. Thus we have
n ~ 1/Ao 3 ~ R 9/4 // 3 . (32.9)
The total number N of degrees of freedom is obtained by multiplying n
by the volume of the region of turbulent flow, which is of the order of Z 3 :f
N  R9/4. (32.10)
Finally, let us consider the properties of the flow in regions whose dimen
sion A is small compared with Ao. In such regions the flow is regular and its
velocity varies smoothly. Hence we can expand v x in a series of powers of
A and, retaining only the first term, obtain v x = constant x A. The order of
magnitude of the constant is v x /Ao, since for A ~ Ao we must have v x ~ v x •
Substituting (32.6) and (32.7)' we find
v x ~ AmR*A//. (32.11)
This formula may also be obtained directly by equating two expressions for
the energy dissipation e: the expression (Am) 3 // (31.1), which determines e
in terms of quantities characterising the large eddies, and the expression
v(w A /A) 2 , which determines e in terms of the velocity gradient (~ v x [\)
for the eddies in which the energy dissipation actually occurs.
PROBLEM
Two fluid particles are at a small distance \ (^> A ) apart. Determine the order of magnitude
of the time t required for the particles to move apart to a distance ^2 (A x <^ Aj <^ /).
Solution. If A ^> A , we have from dimensional considerations dX/dt <~" (cA)*. Integrating
this and using the fact that Ag ^> X lt we find t ~ (V/e)*.
§33. The velocity correlation
Formula (32.1) determines qualitatively the correlation of velocities in
local turbulence, i.e. the relation between the velocities at two neighbouring
points. Let us now introduce quantities which will serve to characterise this
t Formulae (32.6)(32.10) determine how the corresponding quantities vary with the Reynolds
number. Quantitatively, however, it must be borne in mind that a considerable numerical factor
may actually appear in all these formulae. The number of degrees of freedom, for example, must be
of the order of unity not for R <w 1, but for R ^ R cr . Hence we must write the ratio R/R CT in place
of R in (32.10):
N ~ (R/R cr ) 9/4 .
124 Turbulence §33
correlation quantitatively. 4. These may be, for instance, the components of
the tensor
Boc = (»2i«>u)(»2*«>i*)» (33.1)
where V2 and vi are the fluid velocities at two neighbouring points, and the
bar denotes an average with respect to time.J The radius vector from point 1
to point 2 will be denoted by r; we suppose its magnitude r small compared
with / (but not necessarily large compared with the internal scale of turbulence
Ao).
Since local turbulence is isotropic, the tensor J5^ cannot depend on any
direction in space. The only vector that can appear in the expression for Bat
is the radius vector r. In other words, Bm can contain, apart from the absolute
magnitude r of r, only the unit tensor 8^ and the unit vector n in the direction
of r. The most general form of such a tensor of rank two is
B ik = A(r)h ik + B{r)nin k . (33.2)
We take the coordinate axes so that one of them is in the direction of n,
denoting the velocity component along this axis by v r and the component
perpendicular to n by vt. The component B rr is then the mean square
relative velocity of two neighbouring fluid particles along the line joining
them. Similarly, Btt is the mean square transverse velocity of one particle
relative to the other, while B r t is the mean value of the product of these two
velocity components. Since n r = 1, n t = 0, we have from (33.2)
Brr = A + B, B tt = A, Brt=0 (33.3)
Let us now derive a relation between B rr and Btt. To do so, we first
notice that the velocity variation over small distances is mainly due to the
small eddies. The properties of the local turbulence do not depend on the
large eddies that are superposed on it. Hence, to calculate the tensor Buc,
it suffices to take the particular case of completely isotropic and homogeneous
turbulent flow, in which the mean fluid velocity is zero.f f Expanding the
parentheses in (33.1), we have
Bik = VuVi]e + V2iV2k — ViiV2k — VMV2i.
t The results given in this section are due to T. von Karman and L. Howarth (1938) and to A.
N. Kolmogorov (1941). Similar relations for the temperature fluctuations in a nonuniformly heated
turbulent flow are given later (see §54, Problems 3 and 4).
J If there were no correlation between the velocities at the points 1 and 2, the mean values of the
products in (33.1) would reduce to products of the mean value of each factor separately, and would
therefore be zero.
ft Such a flow can be imagined as that of a fluid subjected to strong agitation and then left to itself.
Of course, the flow will certainly decay with time. The averaging in formula (33.1) must then, strictly
speaking, be taken not as an averaging over time but as one over all possible positions of the points
1 and 2 (for a given distance r between them) at a given instant.
§33 The velocity correlation 125
Since the flow is completely homogeneous and isotropic, we have vnvik
= V2tV2k, and vuV2k = vikV2t Thus
Boc = 2vuvijc2vuV2k (33.4)
We differentiate this expression with respect to the coordinates of point 2:
dBijcfdx2k = —2vudv2kl8x2k
By the equation of continuity, however, 8v2kl&X2k = 0, so that dBijddx2k = 0.
Since But is a function only of the components xt = X2i — xu of the vector
r, differentiation with respect to X2k is equivalent to differentiation with
respect to Xk> Substituting (33.2), we have after a simple calculation
A' + B' \2B\r = 0, the prime denoting differentiation with respect to r.
Substituting (33.3), we can write this as B' ' rr + 2(B rr — Btt)jr = 0, whence
we have finally the general relation between B rr and Btt:
2rB tt = d(r*Brr)ldr. (33.5)
At distances r large compared with Ao, the velocity difference is propor
tional to r*, according to (32.1). The components of the tensor Bue for such
r are therefore proportional to r f . Substituting in (33.5) B rr = constant xr f ,
Btt = constant xr s , we obtain the simple relation
Btt = m r . (33.6)
For distances r small compared with Ao, the velocity difference is propor
tional to r, and therefore B rr and Btt are proportional to r 2 . Formula (33.5)
then gives the relation
Btt = 2B rr . (33.7)
At these distances (r <t Ao), Btt and B rr can also be separately expressed in
terms of the mean energy dissipation e. We write B rr = ar 2 , where a is.
constant, and combine (33.2), (33.3), (33.4), obtaining
»u»2* = viiOus ar 2 8 ijc +%ar 2 nink.
Differentiating this relation, we find
dvu dv2i _ dvu di)2i
dxu d%2i dx±i &X2i
Since this holds for arbitrarily small r, we can put xu = X2t, whence
foi\ .,  foi dvi
I I = 15a, = 0.
\ dxi / 8xi dxi
According to the general formula (16.3), however, we have for the mean
126 Turbulence §33
energy dissipation
n / dvt dvi \ 2 [7 dv t \2 dvi dvi 1
e = \v[ + = v][ + = 15av,
\ dxi dxi / L\ dxi 1 Bxi dxi J
whence a = ej!5v. We therefore obtain the following relations giving B rr
and B u in terms of the mean energy dissipation :f
B t = rW 2 M Brr = tW 2 /v. (33.8)
We may also discuss the triple correlation
Buci = {v2i vu)(vz k v lk )(v 2 ivu). (33.9)
We shall again suppose that the flow is completely homogeneous and iso
tropic. Let us first consider the auxiliary tensor vuvi k V2i. This tensor is
symmetrical in the suffixes i and k, and by virtue of the isotropy it must,
like Btk, be expressible in terms of «j and S^. The most general form of such
a tensor is
«ii*>i*0» = C(r)8 ik ni+D(r)(8nn k +S k flii) + F(r)nin k ni. (33.10)
Differentiating with respect to xzu we have by the equation of continuity
dvzi
—(vuV lk V2l) = *>li*>lfc — = 0.
0X21 OX2\
Substituting the expression for vuvi k V2i, we have after a simple calculation
(here omitted) two equations:
d[r*(3C+2D + F)]ldr = 0,
C' + 2(C+D)/r = 0.
Integration of the former gives 3C+2D + F = constant/r 2 . For r = the
functions C, D and F must remain finite. We must therefore put the constant
equal to zero, so that 3C+2D+F = 0. From the two equations thus ob
tained we find
D = (C+irC), F = rC'C. (33.11)
We now expand the parentheses in (33.9). It is easy to see that, by virtue of
t It might be thought that a possibility exists in principle of obtaining a universal formula, appli
cable to any turbulent flow, which should give B rr and Btt for all distances r that are small compared
with /. In fact, however, there can be no such formula, as we see from the following argument. The
instantaneous value of (w 2t  — v u ) (v 2 k —vm) might in principle be expressed as a universal function
of the energy dissipation e at the instant considered. When we average these expressions, however,
an important part will be played by the law of variation of e over times of the order of the periods
of the large eddies (of size <*~> I), and this law is different for different flows. The result of the averaging
therefore cannot be universal.
§33 The velocity correlation 127
the isotropy of the flow, the mean values vuVikVu and V2iV2kV2i are zero.
For all three velocities in these products are taken at the same point; the
only tensor in terms of which the tensor ViVtfvi could be expressed is therefore
hoc. It is, however, impossible to construct a symmetrical tensor of rank
three from unit tensors. Such mean values as vuvi k V2i and t>2i^2fc^iz> on
the other hand, are equal in magnitude and opposite in sign, since the vector
ni in (33.10) changes sign when points 1 and 2 are interchanged. The
result is
Bikl = 2(vi i V 1 ] c V2l + V\iV21cVli + VziWcVll).
Substituting (33.10) and (33.11), we have the expression
B m =2(rC + C)(8 ik m + 8 u n k + 8 k im) + 6{rC G)tynm (33.12)
Again taking one of the coordinate axes parallel to n, we obtain the com
ponents of the tensor Bm". B rrr = — 12C, B r tt = —2(C+rC), B rrt = But
= 0. Hence we see that the relation
6B m = d(rBrrr)ldr (33.13)
holds between the nonzero components B r tt and B rrr .
Finally, it is also possible to find a relation between the components of
the tensors Bik and Bm. To do so, we calculate the derivative d(vuV2k)ldt,
recalling that a completely homogeneous and isotropic flow necessarily
decays with time. Expressing the derivatives dvu/dt and dvzkjdt by means
of the NavierStokes equation, we obtain
d d d 8 /PiV2k\
~iVuV2k) =  (VliVuV2k)  (VuV2kV2l) ~ ~ I 
at dxu 0x21 ox\% \ a I
d I P2V\% \
 +vAl(ai#>2fc) + vA2(t>li«>2&).
OX2k \ P '
In using the properties of homogeneity and isotropy, it must be borne in mind
that the sign of r changes when the points 1 and 2 are interchanged, and
therefore the sign of the (first) space derivatives must be changed. The first
two terms are therefore equal, and so are the last two terms. The third and
fourth terms are zero. For, by virtue of the isotropy, the mean value p\V2k
must be of the form f(r)rik. The divergence d(piV2k)I^X2k = pi 8®2kldx2k is
zero. But the only centrally symmetric vector whose divergence is every
where zero is a constant times (l[r 2 )nk. Such a vector would become infinite
for r = 0, which is impossible. The constant must therefore be zero.
Thus
— (viiV 2 k) =  2 — (vuviMk) + 2v/\iv u v2k (33. 14)
ot dxu
128 Turbulence §34
Here we must substitute, in accordance with the formulae derived above,
VliV2k = ^2i^2& — iBflfc,
(33.15)
— Ts(rB rrr ' — Brr^fiitijcni.
In the former expression we replace V2iV2k by iv 2 Sue, using the complete
homogeneity and isotropy of the flow:
»u*>2* = i« a 8tt&B tt . (33.16)
The time derivative of the kinetic energy per unit mass \v 2 is just the energy
dissipation e; hence d(%v 2 )ldt = §e. A simple, though lengthy, calcula
tion gives the equation
_2
1 dB rr 1 d^Brrr) v d [ dB^
2 dt ~ 6r* dr r* Hr
Kr) (33  17)
Since r is supposed small, we can with sufficient accuracy put r = on the
lefthand side, i.e. neglect 8B rr /dt in comparison with e. Multiplying the
resulting equation by r 4 , integrating over r, and using the fact that the cor
relation functions vanish for r = 0, we obtain the following relation between
Brr and Brrr'
Brrr= Ur + 6vdBrr/dr. (33.18)
The relation (33.18), like (33.13), holds for r either greater or less than Ao.
For r > Ao, the viscosity term is small, and we have simply
Brrr = fer. (33.19)
If r <^ Ao, we can substitute the expression (33.8) for B rr in (33.18), obtaining
Brrr = 0; this is because B rr r in this case must be of the third order in r,
and so the firstorder terms must cancel.f
§34. The turbulent region and the phenomenon of separation
Turbulent flow is in general rotational. However, the distribution of the
vorticity o>( = curl v) in the fluid has certain peculiarities in turbulent flow
(for very large R): in "steady" turbulent flow past bodies, the whole volume
of the fluid can usually be divided into two separate regions. In one of these
the flow is rotational, while in the other the vorticity is zero, and we have
t The ratio  B rrr lB„\ must have constant values in the ranges /^> r ^> A and r <^ A . The ex
perimental results show that in fact this quantity is approximately constant for all r, being about 04.
§34 The turbulent region and the phenomenon of separation 129
potential flow. Thus the vorticity is nonzero only in a part of the fluid
(though not in general only in a finite part).
That such a limited region of rotational flow can exist is a consequence of
the fact that turbulent flow may be regarded as the motion of an ideal fluid,
satisfying Euler's equations.! We have seen (§8) that, for the motion of an
ideal fluid, the law of conservation of circulation holds. In particular, if
at any point on a streamline co = 0, then the same is true at every point
on that streamline. Conversely, if at any point on a streamline co =£ 0,
then co does not vanish anywhere on the streamline. Hence it is clear that
the existence of limited regions of rotational and irrotational flow is compatible
with the equations of motion if the region of rotational flow is such that
the streamlines within it do not penetrate into the region outside it. Such a
distribution of co will be stable, and the vorticity will remain zero beyond
the surface of separation.
One of the properties of the region of rotational turbulent flow is that the
exchange of fluid between this region and the surrounding space can occur in
only one direction. The fluid can enter this region from the region of potential
flow, but can never leave it.
We should emphasise that the arguments given here cannot, of course, be
regarded as affording a rigorous proof of the statements made. However,
the existence of limited regions of rotational turbulent flow seems to be
confirmed by experiment.
The flow is turbulent both in the rotational and in the irrotational region.
The nature of the turbulence, however, is totally different in the two regions.
To elucidate the reason for this difference, we may point out the following
general property of potential flow, which obeys Laplace's equation A<£ = 0.
Let us suppose that the flow is periodic in the ryplane, so that <f> involves
x and y through a factor of the form e iJc i x+ik 2y. Then
320/0*2+ #ty/0y* = (k 1 2 + k 2 2 )<f> = &<(>,
and, since the sum of the second derivatives must be zero, the second deriva
tive of cf> with respect to z must equal <f> multiplied by a positive coefficient:
d 2 cf>l8z 2 =k 2 cf>. The dependence of <f> on z is then given by a damping factor
of the form e~ kz for 2 > (the unlimited increase given by e kz is clearly
impossible). Thus, if the potential flow is periodic in some plane, it must be
damped in the direction perpendicular to that plane. Moreover, the greater
k\ and &2 (i.e. the smaller the period of the flow in the ryplane), the more
rapidly the flow is damped along the sraxis. All these arguments remain
qualitatively valid in cases where the motion is not strictly periodic, but has
only some periodic quality.
From this the following result is immediately obtained. Outside the region
of rotational flow, the turbulent eddies must be damped, and must be so
f The applicability of these equations to turbulent flow ends at distances of the order of A . The
sharp boundary between rotational and irrotational flow is therefore defined only to within such
distances.
130 Turbulence §35
the more rapidly, the smaller their size. In other words, the small eddies do
not penetrate very far into the region of potential flow. Consequently, only
the largest eddies are important in this region; they are damped at distances
of the order of the (transverse) dimension of the rotational region, which is
just the external scale of turbulence in this case. At distances greater than
this dimension there is practically no turbulence, and the flow may be re
garded as laminar.
We have seen that the energy dissipation in turbulent flow occurs in the
smallest eddies; the large eddies do not involve appreciable dissipation,
which is why Euler's equation is applicable to them. From what has been said
above, we reach the important result that the energy dissipation occurs mainly
in the region of rotational turbulent flow, and hardly at all outside that region.
Bearing in mind all these properties of the rotational and irrotational
turbulent flow, we shall henceforward, for brevity, call the region of rotational
turbulent flow simply the region of turbulent flow or the turbulent region.
In the following sections we shall discuss the form of this region in various
cases.
The turbulent region must be bounded in some direction by part of the
surface of the body past which the flow takes place. The line bounding this
part of the surface is called the line of separation. From it begins the surface
of separation between the turbulent fluid and the remainder. The formation
of a turbulent region in flow past a body is called the phenomenon of separa
tion.
The form of the turbulent region is determined by the properties of the
flow in the main body of the fluid (i.e. not in the immediate neighbourhood
of the surface). A complete theory of turbulence (which does not yet exist)
would have to make it possible, in principle, to determine the form of this
region by using the equations of motion for an ideal fluid, given the position
of the line of separation on the surface of the body. The actual position
of the line of separation, however, is determined by the properties of the flow
in the immediate neighbourhood of the surface (known as the boundary
layer), where the viscosity plays a vital part (see §40).
§35. The turbulent jet
The form of the turbulent region, and some other basic properties of it,
can be established in certain cases by simple similarity arguments. These
cases include, among others, various kinds of free turbulent jet in a space
filled with fluid (L. Prandtl, 1925).
As a first example, let us consider the turbulent region formed when a
flow is "separated" at an angle formed by two infinite intersecting planes
(shown in crosssection in Fig. 16). For laminar flow (Fig. 3, §10), the flow
along one side of the angle (AO, say) would turn smoothly and flow along
the other side away from the angle (OB). In turbulent flow, the pattern is
totally different.
The flow along one side of the angle now does not turn on reaching the
§35 The turbulent jet 131
vertex, but continues in its former direction. A flow appears along the
other side in the direction BO. The two flows "mix" in the turbulent
regionjf the boundaries of this region are shown, dashed, in crosssection
in Fig. 16. The origin of this region can be seen as follows. Let us imagine
a flow in which a uniform stream along AO continues in the same direction,
occupying the whole space above the plane AO and its continuation into the
fluid to the right, while the fluid below this plane is at rest. In other words,
we have a surface of separation (the plane AO produced) between fluid moving
with constant velocity and stationary fluid. Such a surface of discontinuity,
however, is unstable, and cannot exist in practice (see §30). This instability
leads to mixing and the formation of a turbulent region. The flow along
BO arises because fluid must enter the turbulent region from below.
777777777777777^^^ V^ *
Let us determine the form of the turbulent region. We take the araxis
in the direction shown in Fig. 16, the origin being at O. We denote by
Yi and Y% the distances from the xzplane to the upper and lower boundaries
of the turbulent region, and require to determine Y± and Y^ as functions of x.
This can easily be done from similarity considerations. Since the planes
are infinite in all directions, there are no constant parameters at our disposal
having the dimensions of length. Hence it follows that Fi, Yi can only be
directly proportional to the distance x:
Yi = a;tanai, Y% = #tana2. (35.1)
The proportionality coefficients are simply numerical constants; we write
them as tan <xi, tan a2, so that <xi and <X2 are the angles between the two
boundaries of the turbulent region and the a?axis. Thus the turbulent region
is bounded by two planes intersecting along the vertex of the angle.
The values of <xi, <X2 depend only on the size of the angle, and not, for
example, on the velocity of the main stream. They cannot be calculated
f We recall that, outside the turbulent region, there is irrotational flow which gradually becomes
laminar as we move away from the boundaries of this region.
132 Turbulence §35
theoretically; the experimental results for flow round a right angle are
ai = 5°, a 2 = 10°.f
The velocities of the flows along the two sides of the angle are not the
same; their ratio is a definite number, again depending only on the size of
the angle. When the angle is not close to it, one of the velocities is considerably
the greater, namely that of the main stream, which is in the same direction
(AO) as the turbulent region. For example, in flow round a right angle, the
velocity along the plane AO is thirty times that along BO.
We may also mention that the difference between the fluid pressures on
the two sides of the turbulent region is very small. For example, in flow round
a right angle it is found that p± — p% = 0003p£/i 2 , where JJ\ is the velocity
of the main stream (along AO), p\ the pressure in that stream, and p2 the
pressure in the stream along BO.
In the limiting case of flow round an angle of 2xr, we have simply the
edge of a plate with fluid moving along both sides. The angle ai + a2 of the
turbulent region is zero, i.e. there is no turbulent region; the velocities of the
flows along the two sides of the plate become equal. As the angle AOB
increases, a point is reached when the plane BO forms the lower boundary of
the turbulent region; the angle AOB is by then obtuse. As the angle increases
further, the turbulent region continues to be bounded by the plane BO on
one side. Here we have simply a separation, with the line of separation along
the vertex of the angle. The angle of the turbulent region remains finite.
As a second example, let us consider the problem of a turbulent jet of
fluid issuing from the end of a narrow tube into an infinite space filled with
the same fluid. The problem of laminar flow in such a "submerged jet" has
been solved in §23. At distances (the only ones we shall consider) large
compared with the dimensions of the mouth of the tube, the jet is axially
symmetrical, whatever the actual shape of the opening.
Let us determine the form of the turbulent region in the jet. We take
the axis of the jet as the #axis, and denote by R the radius of the turbulent
region; we require to determine R as a function of x (which is measured
from the end of the tube). As in the previous example, this function is easily
determined directly from similarity considerations. At distances large
compared with the dimensions of the mouth of the tube, the actual shape and
size of the opening cannot affect the form of the jet. Hence we have at our
disposal no characteristic parameters of the dimensions of length. It therefore
follows as before that R must be proportional to x:
R = x tan a, (35.2)
where the numerical constant tan a is the same for all jets. Thus the turbulent
f Here, and elsewhere, we speak of experimental data on the velocity distribution in a transverse
crosssection of the turbulent jet, reduced by means of calculations (W. Tollmien 1926) based on
the mixinglength theory (see the final note to the present section). This theory contains an arbitrary
constant, whose value is chosen so as to obtain the best possible agreement with experiment.
§35 The turbulent jet 133
region is a cone; the experimental value of the angle 2a is 25 to 30 degrees
(Fig. 17).f
The (time average) velocity distribution in a crosssection of the jet has
the following properties. The flow is principally along the jet. The longitudi
nal velocity component falls off rapidly away from the axis of the jet; it be
comes fwo ("0 being the velocity on the axis) at a distance of only 0351?
from the axis, and at the boundary of the turbulent region it is of the order of
001 mo The transverse velocity component is approximately uniform in order
of magnitude over the crosssection of the turbulent region, and at the
boundary of this region it is about 0025 « , being there directed into the
jet. This transverse component causes a flow into the turbulent region. The
velocity distribution outside the turbulent region (for a given angle a) can
be determined theoretically (see Problem 1).
//
Fig. 17
The velocity in the jet also falls off as we move away from the mouth of
the tube. The law of this decrease is easily found. To do so, we use the
following method. The total flux of momentum through a spherical surface
centred at the tube mouth must be independent of the radius of the surface.
The momentum flux density in the jet is of the order of pu 2 , where u is
of the order of some mean velocity in the jet; this is the only quantity of the
right dimensions that can be formed from the fluid density p, the velocity u,
and the distance x. The area of the part of the jet crosssection where u is
appreciably different from zero is of the order of R 2 . Hence the total momen
tum flux is of the order of pu 2 R 2 . Equating this to a constant and putting
R = constant xx, we obtain
u ~ constant/*, (35.3)
i.e. the velocity diminishes inversely as the distance from the mouth of the
tube.
t Some dependence of the constant a on the initial conditions (velocity profile) in the tube mouth
is observed experimentally. It is reasonable to suppose that this dependence is due to the effect of
the finite dimensions of the opening, an effect which would disappear at greater distances.
134 Turbulence §35
The amount Q of fluid which passes per unit time through a crosssection of
the turbulent region of the jet is of the order of the product of its area ( ~ R 2 )
and the mean velocity u. Substituting, we findf
Q = Bx. (35.4)
Thus the discharge through a crosssection of the turbulent region in
creases with x, i.e. some fluid is, as it were, entrained in the turbulent
region. $ The constant which appears in (35.4) may be determined as follows.
At distances of the order of the dimensions of the tube mouth, Q must be
come the amount Q of fluid emitted from the tube per unit time, which is
fixed for any particular jet. Hence we see that B ~ Qo/a, where a gives the
transverse dimension of the tube mouth (e.g. the radius, if the opening is
circular). Thus we can write
B = cQola, (35.5)
where c is a numerical constant which depends only on the form of the open
ing. If the latter is circular, c is found by experiment to be about 1 5.
The flow in any section of the length of the jet is characterised by the
Reynolds number for that section, defined as uR[v. By virtue of (35.2)
and (35.3), however, the product uR is constant along the jet, so that the
Reynolds number is the same for all such sections. It can be taken, for
instance, as Bjpv. The constant B which appears here is the only parameter
which determines the flow in the jet. When the "strength" Q of the jet
increases (the value of a remaining constant), the Reynolds number Bjpv
eventually reaches a critical value, after which the flow simultaneously
becomes turbulent along the whole length of the jet.ff
t If two variable quantities which vary within wide limits are always of the same order of mag
nitude, then they must be proportional. Hence, in this case (and in similar cases), we can write
precisely Q = constant X a; in place of Q — constant X x.
% The total mass flux through any infinite plane across the jet is infinite, i.e. a jet issuing into an
infinite space carries with it an infinite amount of fluid.
ft In order to make more detailed calculations for various kinds of turbulent flow, it is customary
to employ certain "semiempirical" theories, based on assumptions concerning the dependence of the
turbulent viscosity coefficient on the gradient of the mean velocity. For example, in Prandtl's theory
it is assumed that (for plane flow)
Vturb = l 2 \8uxldy\,
where the dependence of / (called the mixing length) on the coordinates is chosen in accordance
with the results of similarity arguments; for instance, in free turbulent jets we put I = ex, c being an
empirical constant. Such theories usually give good agreement with experiment, and are therefore
useful for interpolatory calculations. However, it is not possible to give universal values to the em
pirical constants which characterise each theory; for example, the value of the ratio of the mixing
length I to the transverse dimension of the turbulent region has to be chosen differently in various
particular cases. It should also be mentioned that good agreement with experimental results can be
obtained with various expressions for the turbulent viscosity.
A more detailed account of these theories is given by L. G. LoItsyanskiI, Aerodynamics of Boundary
Layers (Aerodinamika pogranichnogo sloya), Moscow 1941; G. N. Abramovich, Free Turbulent Jets
of Liquids and Gases (Turbulentnye svobodnye strui zhidkostei i gazov), Moscow 1948; H. Schuchting,
Boundary Layer Theory, Pergamon Press, London 1955.
§35 The turbulent jet 135
PROBLEMS
Problem 1. Determine the mean flow in the jet outside the turbulent region.
Solution. We take spherical coordinates r, 6, $, with the polar axis along the axis of the
jet, and the origin at its point of emergence. Because the jet is axially symmetrical, the
component u^ of the mean velocity is zero, while ug and u r are functions only of r and 6.
The same arguments as were used in the problem of the laminar jet (§23) show that ug and
ur must be of the forms ug — f(0)/r, u r — F(0)Jr. Outside the turbulent region we have
potential flow, i.e. curl u = 0, so that 8ur/d9 — 8(rue)/8r = 0. But rug is independent of r,
so that BurJdO = (1/r) dF/dd = 0, whence F = constant = —b, say, or
ur = bjr. (1)
From the equation of continuity,
Id Id
(r^ur) + —— —iug sin0) = 0,
r 2 dr r sin 9 do
we then obtain
constant— b cos 6
J sin0
The constant of integration must be —b if the velocity is not infinite for d — it (it does not
matter that / is infinite for = 0, since the solution in question refers only to the space
outside the turbulent region, whereas B — lies inside that region). Thus
6 (1 + cos 6) b
u e = r7— =  cotJ0. (2)
r sin 6 r
The component of the velocity in the direction of the jet (u x ) and its absolute magnitude are
b b cos 6 b
u x =  = , u = . . (3)
r x r sin$0
The constant b can be related to the constant B in (35.4). Let us consider a segment of the
cone formed by the turbulent region, bounded by two infinitely close crosssections of the
cone. The mass of fluid entering this segment per unit time is dO = —litrp sin a . ugdr
= 2irbp(l +cos a)dr, while from formula (35.4) we have dQ = B dx = B cos a dr. Com
paring the two expressions, we obtain
B cos a
5 = . (4)
27rp(l + cosa)
At the boundary of the turbulent region, the velocity u is directed into this region, making
an angle \(ir—a.) with the positive direction of the araxis. _
Let us compare the mean velocity u x inside the turbulent region (defined as u z =
QjrrpR 2 = Bjirpx tan 2 a) with the velocity (uz) pot at the boundary of the region. Taking the
first equation (3) with 9 = a, we find
(u x ) P otlu x = 4(1  cos a )'
For a = 12°, this ratio is 0011, i.e. the velocity at the boundary of the turbulent region is
small compared with the mean velocity inside the region.
136 Turbulence §36
Problem 2. Determine the law of variation of size and velocity in a submerged turbulent
jet issuing from an infinitely long thin slit.
Solution. By the same reasoning as for the axial jet, we conclude that the turbulent
region is bounded by two planes intersecting along the slit, i.e. the halfwidth of the jet is
Y = x tan a. The momentum flux in the jet (per unit length of the slit) is of the order of
pu 2 Y. The dependence of the mean velocity u on * is therefore given by u = constant/ V*
The discharge through a crosssection of the turbulent region is Q ~ pu Y, whence Q = con
stant X \/x. The experimental data give a value of 25° to 33° for the angle 2a of a plane
parallel jet (cf. the third footnote to this section).
§36. The turbulent wake
For Reynolds numbers considerably above the critical value, in flow past
a solid body, a long region of turbulent flow is formed behind the body. This
is called the turbulent wake. At distances large compared with the dimension
of the body, simple arguments enable us to determine the form of this wake
and the way in which the fluid velocity decreases there (L. Prandtl, 1926).
As in the investigation of the laminar wake in §21, we denote by U the
velocity of the incident stream, and take the direction of U as the #axis.
The fluid velocity at any point, averaged over the turbulent fluctuations, is
written as U+u. Denoting by a some mean width of the wake, we shall
find a as a function of x. If there is no lift, then at large distances from the
body the wake is axially symmetrical and circular in crosssection ; in this case,
a may be the radius of the wake. If a lift force is present, a direction is
selected in the j#plane, and the wake is not axially symmetrical at any distance
from the body.
The longitudinal fluid velocity component in the wake is of the order of
U, while the transverse component is of the order of some mean value u
of the turbulent velocity. The angle between the streamlines and the *axis
is therefore of the order of u\ U. The boundary of the wake is, as we know, the
boundary beyond which the streamlines of the rotational turbulent motion
cannot pass. Hence it follows that the angle between the boundary of the
wake and the #axis is also of the order of ujU. This means that we can write
dajdx ~ uJU. (36.1)
Next we use formulae (21.1), (21.2), which determine the forces on the
body in terms of integrals of the fluid velocity in the wake (the velocity now
being interpreted as its mean value). The region of integration in these
integrals is of the order of a 2 . Hence an estimate of the integral gives
F ~ pUua 2 , where F is of the order of the drag or the lift. Thus
u ~ F/pUa*. (36.2)
Substituting in (36.1), we find dajdx ~ F/pU 2 a 2 , from which we have by
integration
a ~ {Fx\pTJ 2 )K (36.3)
Thus the width of the wake increases as the cube root of the distance from
§37 Zhukovskii's theorem 137
the body. For the velocity u, we have from (36.2) and (36.3)
u ~ (FU/px*)*, (36.4)
i.e. the mean fluid velocity in the wake is inversely proportional to xK
The flow in any crosssection of the wake is characterised by the Reynolds
number R ~ aujv. Substituting (36.2) and (36.3), we obtain
R ~ FjvpUa ~ {pifPUxiPf.
We see that this number is not constant along the wake, unlike what we found
for the turbulent jet. At sufficiently large distances from the body, R becomes
so small that the flow in the wake is no longer turbulent. Beyond this point
we have the laminar wake, whose properties have been investigated in §21.
In §21, Problem 2, formulae have been obtained which describe the flow
outside the wake and far from the body. These formulae hold for flow
outside the turbulent wake as well as outside the laminar wake.
We may mention here some general properties of the velocity distribution
round the body. Both inside and outside the turbulent wake, the velocity
(by which we always mean u) decreases away from the body. However, the
longitudinal velocity u x falls off more rapidly (~ \\x 2 ) outside the wake
than inside it. Far from the body, therefore, we may suppose u x to be
zero outside the wake. We may say that u x falls from some maximum value
on the axis of the wake to zero at the boundary of the wake. The transverse
components %, u z at the boundary are of the same order of magnitude as
they are inside the wake, diminishing rapidly as we move away from the
wake at a given distance from the body.
§37. Zhukovskii's theorem
The velocity distribution round a body, described at the end of the last
section, does not hold for exceptional cases where the thickness of the wake
formed behind the body is very small compared with its width. A wake
of this kind is formed in flow past bodies whose thickness (in the ^direction)
is small compared with their width (in the ^direction) ; the length (in the
direction of flow, the ^direction) may be of any magnitude. That is, we are
considering flow past bodies whose crosssection transverse to the flow is
very elongated. These bodies include, in particular, wings, i.e. bodies whose
width, or span, is large in comparison with their other dimensions.
It is clear that, in such a case, there is no reason why the velocity com
ponent u y perpendicular to the plane of the turbulent wake should fall off
appreciably at distances of the order of the thickness of the wake. On the
contrary, this component will now be of the same order of magnitude inside
the wake and at considerable distances from it, of the order of the span.
Here, of course, we assume that the lift is not zero, since otherwise the trans
verse velocity practically vanishes.
138 Turbulence §37
Let us consider the vertical lift force F y resulting from such a flow.
According to formula (21.2), it it given by the integral
F y =  P UJju y dydz, (37.1)
where, on account of the nature of the distribution of u y , the integration
must now be taken over the whole transverse plane. Furthermore, since the
thickness of the wake (in the ^direction) is small, while the velocity u y
inside the wake is not large compared with its value outside, we can with
sufficient accuracy take the integration over y to be over the region outside
the wake, writing
Vx
J u y Ay « J u y dy+ \u y dy,
—oo y t —oo
where y 1 and y 2 are the coordinates of the boundaries of the wake (Fig. 18).
£ !/
\ 7/2
\ /I
Fig. 18
Outside the wake, however, we have potential flow, and u y = d<f>/dy;
bearing in mind that <f> = at infinity, we therefore obtain
j Uydy = <f> 2 <f>i,
where $2 and $1 are the values of the potential on the two sides of the wake.
We may say that <f>2<f>i is the discontinuity of the potential at the surface of
discontinuity which may be substituted for a thin wake. The derivative
Uy = dtfajdy must remain continuous. A discontinuity in the velocity com
ponent normal to the surface of the wake would mean that some quantity of
fluid flows into the wake; in the approximation in which the thickness of the
wake is neglected, however, this inflow must be zero. Thus we replace the
wake by a surface of tangential discontinuity. Next, in the same approxima
tion, the pressure also must be continuous at the wake. Since the variation
of the pressure is given in the first approximation, according to Bernoulli's
equation, by pUu x = pU d<j>Jdx, it follows that the derivative dtjyjdx must
also be continuous. The derivative d(j>jdz (the velocity along the wing) is
in general discontinuous, however.
§37 ZhukovsMVs theorem 139
Since the derivative d<f>jdx is continuous, the discontinuity fa<f>i depends
only on z, and not on the coordinate x along the wake. Thus we have
the following formula for the lift:
F y = puffatoP* ( 37  2 )
The integration over z may be taken over the width of the wake (of course,
<j>2 — (f)i = outside the wake).
This formula can be put in a somewhat different form. To do so, we notice
that, using wellknown properties of an integral of the gradient of a scalar,
we can write the difference fafa as a contour integral
& grade/) dl = <j> (u y dy+u x dx),
taken along a contour which starts from the point yi, encircles the body, and
ends at the point y 2 , thus passing at every point through the region of potential
flow. Since the wake is thin we can, without changing the integral except by
quantities of higher order, close this contour by means of the short segment
from y% to y\. Denoting by T the velocity circulation round the closed
contour C enclosing the body (Fig. 18), we have
T = <JJu.dl = ^2^1, (37.3)
and for the lift force the formulaf
F y = pUJTdz. (37.4)
The relation between the lift and the circulation given by this formula
constitutes Zhukovskii's theorem, first derived by N. E. Zhukovskii in 19064
PROBLEMS
Problem 1 . Determine the manner of widening of the turbulent wake formed in transverse
flow past a cylinder of infinite length.
Solution. The drag/x per unit length of the cylinder is of the order of pUu Y. Combining
this with the relation (36.1), we find the width Y of the wake to be
Y = AVixfcJpU*), (1)
where A is a constant. The mean velocity u in the wake falls off in accordance with
u ~ Vifxlpx). The Reynolds number R ~ Yufv ~f x }pUv is independent of x, and there
is therefore no laminar wake.
t The sign of the velocity circulation is always chosen to be that obtained for a counterclockwise
path. The sign in formula (37.3) also depends on the chosen direction of flow. We always suppose
that the flow is in the positive direction of the #axis (from left to right).
% Cf. §46 for the application of this theorem to streamlined wings.
140 Turbulence §38
We may mention that, according to experimental results, the constant coefficient in (1)
is A = 93 ( Y being the halfwidth of the wake ; if Y is taken as the distance at which the
velocity u x falls to half its maximum value (at the centre of the wake), then A — 041).
Problem 2. Determine the flow outside the wake formed in transverse flow past a body
of infinite length.
Solution. Outside the wake we have potential flow; we shall denote the potential by <b
to distinguish it from the angle <fi in the system of cylindrical coordinates which we take,
with the zaxis along the length of the body. As in §21, Problem 2, we conclude that we must
have
cfudf = jgjrad<!>'df = f x /pU,
where now the integration is over the surface of a cylinder of large radius and unit length with
its axis in the adirection, and f x is the drag per unit length of the body. The solution of the
twodimensional Laplace's equation A * = that satisfies this condition is O = (f x J2npU) log r
Next, we have for the lift, by formula (37.2), /„ = pU^O,). The solution of Laplace's
equation that diminishes least rapidly with distance and has a discontinuity on the plane
I = is <D = constant X <f>\ since 4>2~<t>i = 2tt, the constant is fyjlirpU. The flow is given
by the sum of these two solutions, i.e.
The cylindrical components of the velocity u are
u r = d<b\dr = UllirpUr, U<f> = (l/r)a<J>/c# = / y /2*y>tfr. (2)
The velocity u is at a constant angle tan 1 (f y jf x ) to the rdirection.
Problem 3. Determine the manner of bending of the wake behind a body of infinite length
when there is a lift force.
Solution. If there is a lift force, the wake (regarded as a surface of discontinuity) is curved
in the aryplane. The function y = y(x) which determines this is given by the equation
dxJ(ux + U) = djy/tty. Substituting, by (2) of Problem 2, u y « fyllnpVx and neglecting
u x in comparison with U, we obtain
dy/dx = fyllirpUtx,
whence
y = constant (f y /27rpU 2 ) log x.
§38. Isotropic turbulence
We have already mentioned in §33 the particular case of turbulent flow
that is completely homogeneous and isotropic, the mean velocity being zero
throughout the fluid. Such a flow may be imagined as that of a fluid which is
vigorously stirred and then left to itself. The motion decays with time, of
course.
The further investigation of isotropic turbulence, and in particular the
determination of the manner of its decay with time, is based on a conservation
law first derived by L. G. Loitsyanskii (1939). This law, which holds only
for isotropic turbulence, is a consequence of the general law of conservation of
angular momentum, and may be derived as follows.
§38 Isotropic turbulence 141
Let us isolate some fairly large volume in an unbounded fluid, and consider
the total fluid angular momentum M contained in this volume. M has some
random value, which is not in general zero. On account of the interaction with
the surrounding regions, M does not remain strictly constant. However,
since the interaction is a surface effect, it is clear that the time T during which
M varies appreciably must increase with the dimension L of the volume
selected. The time T and the dimension L may be arbitrarily large, and in
this sense the angular momentum M is conserved.
For convenience in what follows, we suppose that the chosen volume of
fluid is enclosed in a vessel with fixed solid walls; it is evident that the boun
dary conditions at the surface of a very large volume cannot have any effect
on the volume properties of the flow, in which we are interested.
According to the general definition, the tensor Mm, which is the total
angular momentum, is equal to the integral
p J (xiV k  x k Vi)dV
taken over the whole volume. We transform this integral as follows :
f x k VidV = f (xiX k vi)dV f xtxt—dV f XiV k dV.
J J dxi J oxi J
The first integral on the righthand side, on being converted into a surface
integral, is seen to be zero, since the normal velocity component at the walls
bounding the fluid is zero, so that vjtdfy = vnd/ = 0. The second integral
is zero if the fluid is incompressible (div v = 0). Thus
and we can write
j x k VidV = — J xivjcdV,
Mm = 2/> J XiVkdV.
The sum of the squared components of Ma is equal to twice the squared
absolute magnitude of the angular momentum vector
M = p j r xvdF.
We therefore have
Af2 = 2/> 2 [ J* XiV k dVf.
The squared integral can be written as a double integral:
M* = 2p2 J* J* Xix'iVw'kdVdV.
142 Turbulence §38
Finally, we notice that this expression may be rewritten
m = P 2 // (xix't) 2 Vkv'kdVdV; (38.1)
the integrals containing the squares xt 2 and x\ 2 vanish, since
// x'^v k v' k dVAV' = J x'WkdV j v k dV, and jv k dV =
because the total linear momentum of an incompressible fluid in a fixed
vessel is zero.
The factor v k v' k = vv' in the integrand of (38.1 is the scalar product
of the velocities at two points having coordinates ^)and x' k , at a distance
r = VK***'*) 2 ] apart. We average this product over all positions of
the points x k and x' k (for given r) in the volume concerned: this averaging
is the same as the one used in §33 in defining the correlation functions. Since
the flow is isotropic, the quantity v^v' is a function of r only. It falls off
rapidly with increasing r, since the velocities of the turbulent flow at two
points a great distance apart may be supposed statistically independent:
the mean value of their product then reduces to the product of the mean values
of the individual velocities, which is zero (the mean velocity being everywhere
zero in the flow under consideration).
Effecting this averaging under the integral sign in (38.1), we find
= P 2 jfdV, where / =  JvV r 2 dV. (38.2)
M 2
The integrand in / diminishes rapidly with increasing r, so that the integral
converges; this means that, as the dimension L of the region tends to infinity,
/ tends to a finite limit. Since the flow is homogeneous,! the quantity /
is constant everywhere in the fluid, and we can write simply M 2 = p 2 /V.
We may point out that the angular momentum is thus found to increase as the
square root of the volume of moving fluid, and not proportionally to the
volume. This is because the total angular momentum is the sum of a large
number of statistically independent components (the angular momenta of
various small portions of the fluid) whose mean values are not zero.
Thus we conclude that, for isotropic turbulence, the constancy of M implies
the condition
f
vv'r 2 dV = constant. (38.3)
This is LoltsyanskiVs law.%
t Throughout the region, except for a very small part near the surface.
% Doubts have recently been expressed more than once concerning the applicability of the con
servation law (38.3), on account of the behaviour of the velocity correlation at very large distances;
for example, if this correlation does not decrease sufficiently rapidly, the integral (38.3) may diverge.
The whole subject seems to be as yet somewhat unclear.
§38 Isotropic turbulence 143
The integrand in (38.3) is noticeably different from zero in a region whose
dimensions are of the order of the scale / of the turbulence (the volume of
the region ~ Z 3 ), and is there of the order of v 2 l 2 . Hence we have from
(38.3)
v 2 l 5 _ constant. (38.4)
Using this relation, we can determine the manner of the time decay of
isotropic turbulence. To do so, we estimate the time derivative of the kinetic
energy of unit volume of the fluid. On the one hand, it may be written as
being of the order of pv 2 jt. On the other hand, it must equal the energy
dissipated in unit volume per unit time. According to formula (31.1),
pe ~ pv*]l (the characteristic velocity here being v). If the two expressions
are comparable, we find
/  vt. (38.5)
Substituting (38.5) in (38.4), we see that
v = constant/t 5/7 . (38.6)
Thus the velocity in isotropic turbulence decays with time inversely as
* 5/7 . For / we have
/= constant xt 2 ", (38.7)
i.e. the external scale of the turbulence increases as t 211 (A. N. Kolmogorov,
1941).
According to formulae (38.6) and (38.7), the Reynolds number R ~ vl/v
decreases as t~ 3n , and after a sufficient time it becomes so small that the
viscosity begins to be important. The energy dissipation is then determined,
on the one hand, by the usual formula (16.3), which gives
/ dvi dvic \ 2 w 2
* = M T~ + r
\ dxjc dxt 1 I 2
and, on the other hand, by e ~ v 2 /t. Comparing, we obtain
/ ~ VH ( 38  8 )
and then from (38.4) we have
v = constant/^/4. (38.9)
These formulae, which are due to M. D. Millionshchikov (1939), give
the manner of decay of isotropic turbulence in the final period, when the
effect of the viscosity becomes predominant.
Isotropic turbulent flow can be brought about by passing a stream through
a grid having a large number of regularly spaced openings. We denote by U
the velocity of the original flow, taking the #axis in the direction of U, and
the true velocity by U+v, so that v is the velocity of the turbulent flow
in which we are interested. If we introduce a frame of reference moving
144 Turbulence §38
with velocity U, then relative to this frame the fluid executes a turbulent
flow with velocity v. As we move away from the grid, the averaged turbulent
flow (with velocity u = v) decays faster than the fluctuating flow. This is
because the averaged flow has a scale of the order of the dimension a of the
grid openings, and these, as we shall see, are small in comparison with the
scale of the fluctuating flow. Consequently, at sufficiently large distances x
from the grid, the averaged velocity u is almost zero, and the turbulent
velocity v is just the fluctuating velocity. At such distances the turbulence
may be regarded as completely isotropic over regions small compared with *
(though not necessarily small compared with the external scale of the tur
bulence). The time decay of the turbulence in the moving frame of reference
corresponds to a decay with increasing distance from the grid in the original
stationary frame. The manner of this decay is given by the formulae derived
above, in which we need only replace / by xjU. Bearing in mind that, at
distances from the grid of the order of a (the dimension of the openings),
we must have / ~ a, we can rewrite formula (38.7) as / ~ a{xja) 2n . For the
velocity we have by (38.5), v ~ lUjx, whence v ~ U{ajx)$ n .
PROBLEM
Using equa tion (3 3.17), obtain for isotropic turbulence the quantitative law of decay of
the quantities v ri Vr2 in the period when viscosity is important (L. G. LoiTSYANSKii, M. D.
MILLIONSHCHIKOV).
Solution. In this case we can neglect the B rrr term in (33.17), as being of a higher order
in the (small) velocity. Introducing the quantity
2v 8 I 8brr\
 — — (r4— : = o.
r 4 8r \ 8r J
brr = V r iVr2 = %V 2 \B r
(see (33.16)), we obtain for it the equation
8b rr 2v 8 / _8b r
~8t
The solution of this equation that is of interest is
b rr = constant xe rV8 "Vi 5/2 ;
cf. the analogous solution (51.6) of the equation of heat conduction. This gives the asymp
totic form of the function b rr for initial conditions such that b rr is any function which de
creases sufficiently rapidly with increasing r (just as (51.6) gives the asymptotic law of pro
pagation of heat which at the initial instant is concentrated in a small region of space).
CHAPTER IV
BOUNDARY LAYERS
§39. The laminar boundary layer
We have several times mentioned the fact that very large Reynolds numbers
are equivalent to very small viscosities, and consequently a fluid may be
regarded as ideal if R is large. However, this approximation can never be
used when the flow in question occurs near solid walls. The boundary con
ditions for an ideal fluid require only the normal velocity component to vanish;
the component tangential to the surface in general remains finite. For a
viscous fluid, however, the velocity at a solid wall must vanish entirely.
From this we can conclude that, for large Reynolds numbers, the decrease
of the velocity to zero occurs almost exclusively in a thin layer adjoining the
wall. This is called the boundary layer, and is thus characterised by the pres
ence in it of considerable velocity gradients. The flow in the boundary layer
may be either laminar or turbulent. In this section we shall consider the
properties of the laminar boundary layer. The boundary of the layer is not,
of course, sharp ; the transition from the laminar flow in it to the main stream
of fluid is continuous.
The rapid decrease of the velocity in the boundary layer is due ultimately
to the viscosity, which cannot be neglected even if R is large. Mathemati
cally, this appears in the fact that the velocity gradients in the boundary
layer are large, and therefore the viscosity terms in the equations of motion,
which contain space derivatives of the velocity, are large even if v is smalL
The mathematical theory of the boundary layer is due to L. Prandtl.
Let us derive the equations of motion of the fluid in a laminar boundary
layer. For simplicity, we consider twodimensional flow along a plane por
tion of the surface. This plane is taken as the xsrplane, with the #axis in
the direction of flow. The velocity distribution is independent of z, and the
velocity has no ^component.
The exact NavierStokes equations and the equation of continuity are then
dv x dv x I dp / d*v x d*v x \
V x 1 V v = h V\ 1 — , (39.1)
x dx Ijy pSx \ 8x* 8y* J K
8Vy 8Vy \dp (d*Vy 8*Vy\
x dx *ly pSy \ 8x* 8y* I V
8?)r 8v v
_1 + _Z = o. (39.3)
dx dy
145
146 Boundary Layers §39
The flow is supposed steady, and the time derivatives are therefore omitted.
Since the boundary layer is thin, it is clear that the flow in it takes place
mainly parallel to the surface, i.e. the velocity v y is small compared with v x
(as is seen immediately from the equation of continuity).
The velocity varies rapidly along the yaxis, an appreciable change in it
occurring at distances of the order of the thickness S of the boundary layer.
Along the #axis, on the other hand, the velocity varies slowly, an appreciable
change in it occurring only over distances of the order of a length / charac
teristic of the problem (the dimension of the body, say). Hence the yderiva
tives of the velocity are large in comparison with the ^derivatives. It follows
that, in equation (39.1), the derivative 8 2 v x /dx 2 may be neglected in compari
son with d 2 v x \dy 2 \ comparing (39.1) with (39.2), we see that the derivative
dp I By is small in comparison with dpjdx (the ratio being of the same order
as Vyjvx). In the approximation considered we can put simply
tylty = 0, (39.4)
i.e. suppose that there is no transverse pressure gradient in the boundary
layer. In other words the pressure in the boundary layer is equal to the pres
sure p(x) in the main stream, and is a given function of x for the purpose of
solving the boundarylayer problem. In equation (39.1) we can now write,
instead of dpjdx, the total derivative dp(x)/dx; this derivative can be ex
pressed in terms of the velocity U(x) of the main stream. Since we have
potential flow outside the boundary layer, Bernoulli's equation, p + ^pU 2
= constant, holds, whence {\jp)dpjdx =  UdJJjdx.
Thus we obtain the equations of motion in the laminar boundary layer in
the form
dv x 8v x d 2 v x 1 dp
v x H v y v =
8x dy dy 2 p dx
dU
= U ,
dx' (39.5)
dv x dvy
+ —  = 0.
dx dy
It can easily be shown that these equations, though derived for flow along a
plane wall, remain valid in the more general case of any twodimensional flow
(transverse flow past a cylinder of infinite length and arbitrary crosssection).
Here x is the distance measured along the circumference of the crosssection
from some point on it, andy is the distance from the surface.
Let Uo be a velocity characteristic of the problem (for example, the
velocity of the main stream at infinity). Instead of the coordinates x, y
and the velocities v x , v y , we introduce the dimensionless variables x', y',
v' x , v' y :
x = lx', y = ///v/R, v x = Uqv' x , v y = Uov'y/^R (39.6)
§39 The laminar boundary layer 147
(and correspondingly U = UoU'), where R = UqIJv. Then the equations
(39.5) take the form
, fa'x , fa'y 8V, &U'
(39.7)
dv' x fa'y
+ —  = 0.
dx' by'
These equations (and the boundary conditions on them) do not involve the
viscosity. This means that their solutions are independent of the Reynolds
number. Thus we reach the important result that, when the Reynolds number
is changed, the whole flow pattern in the boundary layer simply undergoes
a similarity transformation, longitudinal distances and velocities remaining
unchanged, while transverse distances and velocities vary as lfs/K.
Next, we can say that the dimensionless velocities v' x , v' y obtained by
solving equations (39.7) must be of the order of unity, since they do not
depend on R. The same is true of the thickness 8 of the boundary layer in
terms of the coordinates x' y y'. From formulae (39.6) we can therefore
conclude that
Vy ~ Ob/VR. ( 39 « 8 )
i.e. the ratio of the transverse and longitudinal velocities is inversely propor
tional to \/R, and that
S ~ //VR, (39.9)
i.e. the thickness of the boundary layer diminishes with increasing Reynolds
number as l/y^R
Let us apply the equations for the boundary layer to the case of plane
parallel flow along a flat plate. Let the plane of the plate be the xz halfplane
with x > (the leading edge of the plate thus being the line * = 0). We
suppose the plate to extend indefinitely in the positive ^direction. The
velocity of the main stream in this case is evidently constant (U = constant).
The equations (39.5) become
fax fax &*>x fax , fay n mim
v x — + v y = v — , + — = 0. (39.10)
dx dy By* dx dy
The boundary conditions at the surface of the plate are that both velocity
components should vanish: v x = v y = for y = 0, x ^ 0. As we move
away from the plate, the velocity must approach asymptotically the velocity
U of the incident flow, i.e. v x = U for y > ± oo. In the solution of the
equations for the boundary layer, as we have seen, v x jU and v y <s/{HUv) can
be functions only of x' = xjl and y' = y\/(Ullv). In the problem under
consideration, however, the plate is infinite in extent and there are no charac
teristic lengths /. Hence v x jU can depend only on a combination of x' and y'
148 Boundary Layers §39
which does not involve /, namely y'j\/x' = y\Z(Ufvx). Similarly, the product
v'yV x ' must be a function of y /\/x'. Thus we can seek a solution in the form
** = Uf[yV(Ulvx)l v y = ViUvlxfilyy/iU/vx)], (39.11)
where/ and /i are some dimensionless functions. Using the second equation
(39.10), we can express /i in terms of/. The problem thus reduces to the
determination of a single function / of a single variable  = y\/{UJvx).\
In what follows we shall be interested only in the distribution of the
longitudinal velocity v x (since v y is small). We can draw an important
conclusion from formula (39.11) without even determining the function/.
The velocity v x increases from zero at the surface of the plate to a definite
fraction of Ufor a given value of the argument of/, i.e. for y y/{ U/vx) = any
given constant. Hence we can conclude that the thickness of the boundary
layer in flow along a plate is given in order of magnitude by
8 ~ V(vx/U). (39.12)
Thus, as we move away from the edge of the plate, 8 increases as the square
root of the distance from the edge.
The function / can be determined by numerical integration. A graph of
this function is shown in Fig. 19. We see that / tends very rapidly to its
limiting value of unity. J
The frictional force on unit area of the surface of the plate is
o xy = 7]{dv x jdy) y= Q.
A numerical calculation gives
o xy = 0332V(^C/ 3 /^). (39.13)
If the plate is of length / (in the xdirection), then the total frictional force
on it per unit length in the ^direction is
i
F = 2 a X ydx.
o
The factor 2 is due to the fact that the plate has two sides exposed to the
t It is easily shown that, if the function <£(£) is such that/(£) = <f>'(£), then/ 1 (f) = 4(£<£'— <f>),
while (f> satisfies the equation <fxf>"+2<l>'" = 0, with the boundary conditions <j> = </>' = for £ = 0,
</>' = 1 for £ = oo.
X The "displacement thickness" 8*, sometimes used to characterise the thickness of the boundary
layer, is defined by
00
j (Uv x )dy = US*.
o
It is equal to 1*72 \/(vx/U).
§39
The laminar boundary layer
149
fluid. Substituting (39.13), we have
F = 1328 VMC/ 8 ) ( 39  14 )
(H. Blasius, 1908). We may point out that the frictional force is proportional
to the f power of the velocity of the main stream. Formula (39.14) can be
applied only to fairly long plates, for which the Reynolds number Uljv is
fairly large. The force is customarily expressed in terms of the drag coefficient,
defined as the dimensionless ratio
C = FfoW.Zl. (39.15)
By (39.14), this quantity, for laminar flow along a plate, is inversely pro
portional to the square root of the Reynolds number:
C = 1328/VR. C 39  1 ^)
The quantitative formulae obtained above relate, of course, only to flow
along a flat plate. The qualitative results, however, such as (39.8) and (39.9),
hold for flow past bodies of any shape; in such cases / is the dimension of
the body in the direction of flow.
10
08
06
04
02
3 4 5
Fig. 19
We may make special mention of two cases of the boundary layer. If we
have a plane disk, of large radius, rotating in the fluid about an axis perpen
dicular to its plane, then to estimate the thickness of the boundary layer we
must replace U in (39.12) by Q.x, where Q is the angular velocity of rotation.
We then find
S ~ \/(v/Q). (39.17)
We see that the thickness of the boundary layer may be regarded as a constant
over the surface of the disk, in accordance with the exact solution of this
problem obtained in §23. The magnitude of the frictional forces on the disk,
as obtained from the equations for the boundary layer, is of course (23.4),
since this formula is exact and therefore holds for laminar flow with any
value of R.
150 Boundary Layers §39
Finally, let us consider the laminar boundary layer formed at the walls of
a pipe near the point of entry of fluid. The fluid usually enters the pipe with
a velocity distribution which is almost constant over the crosssection, and
the velocity falls to zero entirely within the boundary layer. As we move
away from the entrance to the pipe, the fluid layers nearer the axis are re
tarded. Since the mass of fluid that passes each crosssection is the same, the
inner part of the stream, where the velocity is still uniform, must be
accelerated as its diameter is reduced. This continues until a Poiseuille
velocity distribution is asymptotically reached; this distribution is thus found
only at some distance from the entrance to the pipe. It is easy to determine
the order of magnitude of the length / of the "inlet section". It is given by the
fact that, at a distance / from the entrance, the thickness of the boundary
layer is of the same order of magnitude as the radius a of the pipe, so that the
boundary layer fills almost the whole crosssection. Putting in (39.12)
x ~ / and S ~ a, we obtain
/ ~ a 2 Ujv ~ aR. (39.18)
Thus the length of the inlet section is proportional to the Reynolds number, f
PROBLEMS
Problem 1 . Determine the thickness of the boundary layer near a stagnation point (see §10).
Solution. Near the stagnation point the fluid velocity (outside the boundary layer) is
proportional to the distance x from that point, so that we can put U — ex. By estimating the
magnitudes of the terms in the equations (39.5) we find S ~ \/(v/c). Thus the thickness of
the boundary layer near the stagnation point is finite (and, in particular, does not vanish at
the stagnation point itself).
Problem 2. Determine the flow in the boundary layer in a converging channel between
two nonparallel planes (K. Pohlhausen, 1921).
Solution. Considering the boundary layer along one of the planes, we measure the co
ordinate x along that plane from the point O (Fig. 8, §23). For an ideal fluid we should have
the velocity U = Qfoucp ; the corresponding pressure gradient is, by Bernoulli's equation, given
by
Id* d O 2
T"= (C/2)= *
p dx dx a 2 # 3 /o 2
It is easy to see that v x and v y must be sought in the form
v x = (Qlpooc)f(y/x), v y = {Qlp«x)fi(ylx).
From the equation of continuity we obtain f t = (y/x)f, and the first equation (39.5) then gives
for the function /
(W0/"=l/2,
where the prime denotes the differentiation of/ with respect to its argument £ = y/x. The
f We shall not discuss the theory of the boundary layer for a compressible fluid, which is, of course,
considerably more complicated than that for an incompressible fluid. An account of this theory may
be found in: N. E. Kochin, I. A. Kibel' and N. V. Roze, Theoretical Hydromechanics (Teoreticheskaya
gidromekhanika), Part 2, 3rd ed., Chapter II, §§35, 36, Moscow 1948; H. Schlichting, Boundary
Layer Theory, Pergamon Press, London 1955; L. Howarth ed., Modern Developments in Fluid
Dynamics: High Speed Flow, vol. 1, Oxford 1953.
§40 Flow near the line of separation 151
boundary conditions are /(0) = 0, /(co) = 1 (since we must have (v x )y=f> = 0, (v x )y=(o
= Qlpcuc). A first integral of the equation is
(vap/20)/' 2 =/£/3+ constant.
Since /tends to unity as y *■ oo, we see that/' tends to a definite limit, which can only be
zero. The constant being thereby determined, we find
(mp/2£)/'2 = _ K /l)2( /+2 ).
Since the righthand side is always negative for ^ / < 1, we must have Q < 0. That is,
a boundary layer of the type in question is formed only by flow in a converging channel
(and only at large Reynolds numbers R = \Q\/vp), and not by flow in a diverging channel, in
accordance with the results of §23. Integrating again, we have finally
/ = 3 tanh2[log( V2 + VV + fl/W 2 ")]  2
§40. Flow near the line of separation
In describing the line of separation (§34) we have already mentioned that
the actual position of this line on the surface of the body is determined by the
properties of the flow in the boundary layer. We shall see below that, from a
mathematical point of view, the line of separation is a line whose points
are singular points of the solutions of (PrandtPs) equations of motion in the
boundary layers. The problem is to determine the properties of these solu
tions near such a line of singularities.!
We know already that, from the line of separation, there begins a surface
which extends into the fluid and marks off the region of turbulent flow. The
flow is rotational throughout the turbulent region, whereas in the absence of
separation it would be rotational only in the boundary layer, where the vis
cosity is important; the vorticity would be zero in the main stream. Hence
we can say that separation causes the vorticity to "penetrate" from the boun
dary layer into the fluid. By the conservation of circulation, however,
this "penetration" can occur only by the direct mixing of fluid moving near
the surface (in the boundary layer) with the main stream. In other words, the
flow in the boundary layer must be separated from the surface of the body, the
streamlines consequently leaving the surface layer and entering the interior
of the fluid. This phenomenon is therefore called separation or separation of
the boundary layer.
The equations of motion in the boundary layer lead, as we have seen, to the
result that the tangential velocity component (v x ) in the boundary layer is
large compared with the component (%) normal to the surface of the body.
This relation between v x and v y derives from our basic assumptions regarding
the nature of the flow in the boundary layer, and must necessarily be found
wherever Prandtl's equations have physically meaningful solutions. Mathe
matically, it is found at all points not lying in the immediate neighbourhood
of singular points. But if v y <^ v x it follows that the fluid moves along the
f The treatment of the problem given here, due to L. D. Landau, is somewhat different from
that usually given.
152 Boundary Layers §40
surface of the body, and moves away from the surface only very slightly,
so that there can be no separation. We therefore reach the conclusion that
separation can occur only on a line whose points are singularities of the
solution of PrandtPs equations.
The nature of these singularities also follows immediately. For, as we
approach the line of separation, the flow deviates from the boundary layer
towards the interior of the fluid. In other words, the normal velocity com
ponent ceases to be small compared with the tangential component, and is
now of at least the same order of magnitude. We have seen (cf. (39.8))
that the ratio v y jv x is of the order of Ijy/R, so that an increase of v y to the
point where v y ~ v x means an increase by a factor of y'R. Hence, for suffi
ciently large Reynolds numbers (which, of course, we are considering)
we may suppose that v y increases by an infinite factor. If we use Prandtl's
equations in dimensionless form (see (39.7)), the situation just described is
formally equivalent to an infinite value of the dimensionless velocity v' y
on the line of separation.
In order to simplify the subsequent discussion a little, we shall consider
the twodimensional problem of transverse flow past a body of infinite length.
As usual, x is the coordinate along the surface in the direction of flow, while
y is the distance from the surface of the body. Instead of a line of separation,
we now have a point of separation, namely the intersection of the line of
separation with the xyplane; in the coordinates used, this is the point
x = constant s xo, y = 0. Let x < xo be the region in front of the point
of separation.
According to the above results, we have for allf y
v y (xo,y) = oo. (40.1)
In Prandtl's equations, however, v y is a kind of parameter, which is usually
of no interest (on account of its smallness) in investigating the flow in the
boundary layer. Hence it is necessary to ascertain the properties of the func
tion v y near the line of separation.
It is clear from (40.1) that, for x = xo, the derivative dv y J8y also becomes
infinite. From the equation of continuity, dv x \dx\dv y \dy = 0, it then fol
lows that (dv x /8x) x==Xo is infinite, or dxjdv x = 0, where x is regarded as a
function of v x and y. We denote by ©o(y) the value of the function v x (x, y)
for x = xq: vo(y) = v x (xq, y). Near the point of separation, the differences
v x — vo and xq — x are small, and we can expand xq — x in powers of v x — vo
(for a given y). Since (dx/dv x ) v=v<) = 0, the firstorder term in this expansion
must vanish identically, and we have as far as terms of the second order
xox =f(y)(v x v ) 2 , or
v% = My) + *{y)V( x o  x )> (40.2)
f Except y = 0, where we must always have v y — in accordance with the boundary conditions
at the surface of the body.
§40 Flow near the line of separation 153
where a = Ijy/f is some function of y alone. Putting now
foy _ dvx _ afcy)
dy &v 2\/(#o — #)
and integrating, we have for v y
vy = Ky)!V(xox)> (403)
where
is another function of _y.
Next, we use the first equation (39.5):
dv x &v x d*v x 1 dp
a^ + v y = v — — . (40.4)
dx dy dy 2 p dx
The derivative d z v x jdy 2 does not become infinite for x = xo, as we see
from (40.2). The same is true of dp/cbc, which is determined by the flow
outside the boundary layer. Both terms on the lefthand side of equation
(40.4) become infinite, however. In the first approximation we can therefore
write for the region near the point of separation v x dv x jdxhv y dv x jdy = 0.
Substituting dv x jdx = — dvy/dy, we can rewrite this as
8v y dv x d
Vy— = V x 2
 ()o
By dy 8y \ v x I
Since the velocity v x does not in general vanish for x = xo, it follows that
8(vylv x )Jdy = 0, i.e. the ratio v y jv x is independent of y. From (40.2) and
(40.3), we have to within terms of higher order
v% woOOVfao*)
If this is a function of x alone, we must have fi(y) = %Avo(y), where A is a
numerical constant. Thus
Vy . **> , (40.5)
2y/(x — x)
Finally, noticing that a and /S in (40.2) and (40.3) obey the relation a = 2j8',
we obtain en — A dvojdy, so that
v x = v {y) + A(dv ldy) \Z(x  x). (40.6)
Formulae (40.5) and (40.6) determine v x and v y as functions of x near
the point of separation. We see that each can be expanded in this region in
powers of \^{xq — x), the expansion of v y beginning with the —1 power, so
154 Boundary Layers §40
that v y becomes infinite as (#0 — x)~l for x > xq. For x > xq, i.e. beyond
the point of separation, the expansions (40.5) and (40.6) are physically
meaningless, since the square roots become imaginary; this means that the
solutions of Prandtl's equations which give the flow up to the point of
separation cannot meaningfully be continued beyond that point.
From the boundary conditions at the surface of the body, we must always
have v x = v y = for y = 0. We therefore conclude from (40.5) and (40.6)
that
v (0) = 0, (dvoldy)y =0 = 0. (40.7)
Thus we have the important result (due to Prandtl) that, at the point of
separation itself (x = xo, y = 0), not only the velocity v x but also its first
derivative with respect to y is zero.
It must be emphasised that the equation dv x /dy = on the line of separa
tion holds only when v y becomes infinite for that value of x. If the constant
A in (40.5) happens to be zero, so that v v {xq, y) # 00, then the point x = xq,
y = at which the derivative 8v x /8y vanishes would have no other particular
properties, and would not be a point of separation. A can vanish, however,
only by chance, and such an event is therefore unlikely. In practice a point
on the surface of the body at which dv x jdy = is always a point of separation.
If there is no separation at the point x = xo (i.e. if A = 0), then for
x > xq we have (8v x /dy) y =o < 0, i.e. v x becomes negative (of increasing
absolute magnitude) as we move away from the surface, y being still small.
That is, the fluid beyond the point x = xq moves, in the lower parts of the
boundary layer, in the direction opposite to that of the main stream ; there is a
"backflow" of fluid at this point. It must be emphasised that from such
arguments we cannot conclude that there is necessarily a point of separation
where dv x jdy = 0; the whole flow pattern with the "backflow" might lie
(as it does for A = 0) entirely within the boundary layer and not enter the
main stream, whereas it is characteristic of separation that the flow enters
the main body of the fluid.
It has been shown in the previous section that the flow pattern in the boun
dary layer is similar for different Reynolds numbers, and, in particular the
scale in the ^direction remains unchanged. It follows from this that the
value #0 of the coordinate x for which the derivative {dv x jdy) y =o is zero is
the same for all R. Thus we have the important result that the position of
the point of separation on the surface of the body is independent of the
Reynolds number (so long as the boundary layer remains laminar, of course ;
see §45).
Let us also ascertain the properties of the pressure distribution p(x)
near the point of separation. For y = the lefthand side of equation (40.4)
is zero together with v x and v y , and there remains
v(^/^ 2 )h = (Vp)dpldx. (40.8)
It is clear from this that the sign of dpfdx is the same as that of (d 2 v x ]dy z ) y =o.
§40 Flow near the line of separation 155
When (dvxldy)y=o > we can say nothing regarding the sign of the second
derivative. However, since v x is positive and increases away from the wall
(in front of the point of separation), we must always have (d 2 %/#y 2 )2/=o >
at x = xo itself, where dv x /dy = 0. Hence we conclude that
(dp/dx) x = Xa > 0, (40.9)
i.e. the fluid near the point of separation moves from the lower pressure to the
higher pressure. The pressure gradient is related to the gradient of the
velocity U(x) outside the boundary layer by (l/p)dp/d# = — U dU/dx.
Since the positive direction of the axis is the same as the direction of the
main stream, U > 0, and therefore
(dU/dx) x = Xa < 0, (40.10)
i.e. the velocity U decreases in the direction of flow near the point of separa
tion.
From the results obtained above we can deduce that there must be separa
tion somewhere on the surface of the body. For there is on both the front and
the back of the body a point (the stagnation point) at which the fluid velocity
is zero for potential flow of an ideal fluid. Consequently, for some value of x,
the velocity U(x) must begin to decrease, and finally it becomes zero. It is
clear, however, that the fluid moving over the surface of the body is retarded
the more strongly, the closer it is to the surface (i.e. the smaller is y). Hence,
before the velocity U(x) is zero at the outer limit of the boundary layer, the
velocity in the immediate neighbourhood of the surface must be zero. Mathe
matically, this evidently means that the derivative dv x /dy must always vanish
(and therefore there must be separation) for some x less than the value for
which U(x) = 0.
In flow past bodies of any form the calculations can be carried out in an
entirely similar manner, and they lead to the result that the derivatives
dv x [dy, dvz/dy of the two velocity components v x and v z tangential to the
surface of the body vanish on the line of separation (the jyaxis, as before,
is along the normal to the portion of the surface considered).
We may give a simple argument which demonstrates the necessity of separa
tion in cases where the fluid would otherwise have a rapid increase of pressure
(and therefore a rapid decrease in the velocity U) in the direction of its flow
past the body. Over a small distance A.x = x% — xi, let the pressure p
increase rapidly from/>i top2 (j>2 > Pi) Over the same distance A#, the fluid
velocity U outside the boundary layer falls from its initial value U\ to a
considerably smaller value TJi determined by Bernoulli's equation:
i(Ul 2 U2 2 )=(p2pl){p.
Since p is independent of y, the pressure increase p<i —p± is the same at all
distances from the surface. If the pressure gradient dpfdx ~ (j>2—pi)l^x
is sufficiently high, the termvd^^dy 2 involving the viscosity may be omitted
from the equation of motion (40.4) (if, of course, y is not small). Then, to
156 Boundary Layers §41
estimate the change in the velocity v in the boundary layer, we can use
Bernoulli's equation, putting %{v2 2 — vi 2 ) = — (P2—pi)lp, or, from the
equation previously obtained, V2 2 = v± 2 — (U1 2 — U2 2 ). The velocity v\ in
the boundary layer is less than that of the main stream, and we can select
a value of y for which vi 2 < Ui 2 — U2 2 . The velocity V2 is then imaginary,
showing that Prandtl's equations have no physically significant solutions.
In fact, there must be separation in the distance Ax, as a result of which the
pressure gradient is reduced.
An interesting case of the appearance of separation is given by flow at an
angle formed by two intersecting solid surfaces. For laminar potential flow
outside an angle (Fig. 3), the fluid velocity at the vertex of the angle would
become infinite (see §10, Problem 6), increasing in the stream approaching the
vertex and diminishing in the stream leaving the vertex. In reality, the rapid
decrease in velocity (and corresponding increase in pressure) beyond the
vertex would lead to separation, the line of separation being the line of
intersection of the surfaces. The resulting flow pattern is that discussed in
§35.
In laminar flow inside an angle (Fig. 4), the fluid velocity is zero at the
vertex. In this case the velocity diminishes (and the pressure increases) in
the flow approaching the vertex. The result is in general the appearance of
separation, the line of separation being upstream from the vertex of the angle.
PROBLEM
Determine the least possible increase Ap in the pressure which can occur (in the main
stream) over a distance Ax and cause separation.
Solution. Let y be a distance from the surface of the body at which, firstly, Bernoulli's
equation can be applied and, secondly, the squared velocity v 2 (y) in the boundary layer is
less than the change  A[/ 2  in the squared velocity outside that layer. For v(y) we can write,
in order of magnitude, v(y) K y dv/dy ~ Uy/S, where S ^ \/(vlJU) is the width of the
boundary layer and I the dimension of the body. Equating, in order of magnitude, the two
terms on the righthand side of equation (40.4), we find
(l/p)A^/Ax ~ w(y)ly 2 ~ vUjBy.
From the condition
v 2 = \AU 2 \ = {2jp)Ap we have U 2 y 2 l8 2 ~ Apjp.
Eliminating y, we finally obtain
Ap ~ P U 2 (Axll)K
§41. Stability of flow in the laminar boundary layer
Laminar flow in the boundary layer, like any other laminar flow, becomes to
some extent unstable at sufficiently large Reynolds numbers. The manner
of the loss of stability in the boundary layer is similar to that which occurs for
flow in a pipe (§29).
The Reynolds number for flow in the boundary layer varies over the surface
§41 Stability of flow in the laminar boundary layer 157
of the body. For example, in flow along a plate we could define the Reynolds
number as Ra; = Uxjv, where x is the distance from the leading edge of the
plate, and U the fluid velocity outside the boundary layer. A more suitable
definition for the boundary layer, however, is one in which the length para
meter directly characterises the thickness of the layer; such, for instance, is the
"displacement thickness" S* (see the second footnote to §39). We then
have R§* = U8*jv. Since the dependence of the boundarylayer thickness
on the distance x is given by formula (39.12), it is clear that R§* ~ V^t
Because the change of the layer thickness with distance is comparatively
slow, it may be neglected in investigating the stability of flow in a small
portion of the layer, and we may consider a rectilinear twodimensional flow,
with a velocity profile which does not vary along the ^axis.J Then, from a
mathematical point of view, the problem is entirely analogous to that of
the stability of flow between two parallel planes, discussed in §29. The only
difference is in the form of the velocity profile; instead of a symmetrical
profile with v = on both sides, we now have an unsymmetrical profile in
which the velocity changes from zero at the surface of the body to some
given value U, the velocity of the flow outside the boundary layer. The
investigation leads to the following results (Lin, 1945 ; see C. C. Lin, The
Theory of Hydrodynamic Stability, Cambridge 1955).
The form of the limiting curve of stability in the coRplane (see §29)
depends on the form of the velocity profile in the boundary layer. If the
velocity profile has no point of inflexion, and the velocity v x increases
monotonically with the curve v x = v x {y) everywhere convex upwards (Fig.
20a), then the boundary of the stable region is completely similar in form to
that which is obtained for flow in a pipe: there is a minimum value R = R cr
at which amplified perturbations first appear, and for R > oo both branches
of the curve are asymptotic to the axis of abscissae (Fig. 21a). For the velocity
profile which occurs in the boundary layer on a flat plate, the critical Reynolds
number is found by calculation to be R§*,cr ~ 420.ft
A velocity profile of the kind shown in Fig. 20a cannot occur if the fluid
velocity outside the boundary layer decreases downstream. In this case the
velocity profile must have a point of inflexion. For, let us consider a small
portion of the surface, which we may regard as plane, and let x be again the
coordinate in the direction of flow, and y the distance from the wall. From
(40.8) we have
v(d 2 v x lfy 2 )y=o = (llp)dpldx = UdU/dx,
whence we see that, if U decreases downstream (dUjdx < 0), we must have
d^x/dy 2 > near the surface, i.e. the curve v x = Vx(y) is concave upwards.
As y increases, the velocity v x must tend asymptotically to the finite limit U.
t For example, in a laminar boundary layer on a flat plate R^» = 1  72\/R:r.
j In doing so, of course, we pass over the question of the effect which the curvature of the surface
may have on the stability of the boundary layer.
ft For R»* > oo, co tends to zero, on the two branches I and II of the limiting curve, as R^*  * and
R^* 1 / 5 respectively.
158
Boundary Layers
§41
It is then clear from geometrical considerations that the curve must become
convex upwards, and therefore must have a point of inflexion (Fig. 20b).
In this case the form of the curve defining the stable region is slightly changed :
the two branches have different asymptotes for R ► oo, one tending to the
axis of abscissae and the other to a finite nonzero value of o> (Fig. 21b).
The presence of a point of inflexion also reduces considerably the value of R cr .
The fact that the Reynolds number increases along the boundary layer
makes the behaviour of the perturbations as they are carried downstream
somewhat unusual. Let us consider flow along a flat plate, and suppose that a
perturbation of given frequency co occurs at some point in the boundary
layer. Its propagation downstream corresponds to a movement in Fig. 21a
to the right along a horizontal line a> = constant. The perturbation is at
first damped : then, on reaching branch I of the stability curve, it begins to
be amplified. This continues until branch II is reached, whereupon the
perturbation is again damped. The total "amplification coefficient" for the
perturbation during its passage through the region of instability increases
very rapidly as this region moves towards large R (i.e. as the corresponding
horizontal segment between branches I and II moves downwards).
These results, however, do not answer the question whether true absolute
instability occurs in the laminar boundary layer for sufficiently large R —
that is, instability due to the amplification in time of perturbations at a given
point (see §29). As with flow in a pipe, no such investigation has yet been
made.
§42
The logarithmic velocity profile
159
The experimental results for flow along a flat plate show that the point
where turbulence appears in the boundary layerf depends to a considerable
extent on the intensity of the perturbations in the main stream. For marked
perturbations, the boundary layer was observed to become turbulent for
Rs* » 560. As the intensity of the perturbations diminishes, the onset of
turbulence is postponed to higher values of R s «, which seem to tend to a
finite limit of about 3000.
It is possible that the existence of the limit indicates the presence of true
absolute instability for sufficiently high values of R. On the other hand,
it may be that, because of the extremely rapid increase of the "amplification
coefficient" with R, the "displacement" instability of the kind described above
may give the appearance of true instability.
§42. The logarithmic velocity profile
Let us consider planeparallel turbulent flow along an unbounded plane
t Because the Reynolds number varies along the plate, the whole boundary layer does not become
turbulent immediately, but only the part where R^» exceeds a certain value. For a given incident
velocity, this means that turbulence begins at a definite distance from the leading edge; as the velocity
increases, this distance approaches zero.
160 Boundary Layers §42
surface; the term "planeparallel" applies, of course, to the time average of the
flow.f We take the direction of the flow as the araxis, and the plane of the
surface as the xsrplane, so that y is the distance from the surface. The y and
z components of the mean velocity are zero: u x = u, u y = u z = 0. There
is no pressure gradient, and all quantities depend on y only.
We denote by a the frictional force on unit area of the surface ; this force is
clearly in the ^direction. The quantity o is just the momentum trans
mitted by the fluid to the surface per unit time; it is the constant flux of
the ^component of momentum, which is in the negative jydirection, and
gives the amount of momentum transmitted from the layers of fluid remote
from the surface to those nearer it.
The existence of this momentum flux is due, of course, to the presence of a
gradient, in the jydirection, of the mean velocity u. If the fluid moved
with the same velocity at every point, there would be no momentum flux.
The converse problem can also be stated: given some definite value of a,
what must be the motion of a fluid of given density p to give rise to a momen
tum flux a? For large Reynolds numbers, the viscosity v is, as usual, unim
portant; it becomes important only for small distances y (see below). Thus
the value of the velocity gradient dujdy at each point must be determined by
the constant parameters p, a and, of course, the distance y itself. The
dimensions of these quantities are respectively g/cm 3 , g/cm sec 2 and cm.
The dimensions of the derivative dujdy are 1/sec. The only combination
of p, a and y that has the right dimensions is Vi^lpy 2 ) Hence we must have
dujdy = vWrifty. (42.1)
where b is a numerical constant ; b cannot be calculated theoretically, and
must be determined experimentally. It is found to bej
b = 0417. (42.2)
We introduce the more convenient notation v% = ^{crjp), so that
a = /w # 2 . (42.3)
The quantity v% has the dimensions cm/sec and acts as a characteristic velocity
for the turbulent flow considered; then (42.1) becomes dujdy = v%jby,
whence
« = (»#/*)(log^ + c). (42.4)
where c is a constant of integration. To determine this constant we cannot
use the ordinary boundary conditions at the surface, since for y = the first
term in (42.4) becomes infinite. The reason for this is that the above expres
sion is really inapplicable at very small distances from the surface, since the
effect of the viscosity then becomes important, and cannot be neglected.
f The results given in §§4244 are due to T. von Karman and L. Prandtl.
j The value of this constant, and of one in formula (42.8) below, are obtained from measurements
of the velocity distribution near the walls of a pipe in which there is turbulent flow.
§42 The logarithmic velocity profile 161
There are also no conditions at infinity, since for y = oo the expression
(42.4) again becomes infinite. This is because, in the idealised conditions
which we have imposed, the surface is unbounded, and its influence therefore
extends to infinitely great distances.
Before determining the constant c, we may first point out the following
important property of the flow considered: contrary to what usually happens,
it has no characteristic constant parameters of length which might give the
external scale of the turbulence. This scale is therefore determined by the
distance y itself: the scale of turbulent flow at a distance j from the surface is
of the order of y. The fluctuating velocity of the turbulence is of the order of v m .
This also follows at once from dimensional arguments, since v% is the only
quantity having the dimensions of velocity which can be formed from the
quantities or, p, y at our disposal. It should be emphasised that, whereas the
mean velocity decreases with y, the fluctuating velocity remains of the same
order of magnitude at all distances from the surface. This result is in accor
dance with the general rule that the order of magnitude of the fluctuating
velocity is determined by the variation Am of the mean velocity (§31). In
the present case, there is no characteristic length / over which the variation
of the mean velocity could be taken; Am must now be defined, reasonably,
as the change in u when the distance y changes appreciably. According to
(42.4), such a change in y causes a change in the velocity u that is just of the
order of v m .
At sufficiently small distances from the surface, the viscosity of the fluid
begins to be important; we denote the order of magnitude of these distances
by yo, which can be determined as follows. The scale of the turbulence at
these distances is of the order of yo, and the velocity is of the order of v m .
Hence the Reynolds number which characterises the flow at distances of the
order of jo is R ~ v^yolv. The viscosity begins to be important when R
becomes of the order of unity. Hence we find that
yo ~ vjv mt (42.5)
and this determines yo.
At distances from the surface small compared with yo, the flow is deter
mined by ordinary viscous friction. The velocity distribution here can be
obtained directly from the usual formula for viscous friction : a = pv dujdy,
whence
u = ayjpv = v m 2 y/v. (42.6)
Thus, immediately adjoining the wall, there is a thin layer of fluid in which
the mean velocity varies linearly with y; the velocity is small throughout
this layer, varying from zero at the surface itself to values of the order of
v m for y ~ yo. We shall call this layer the viscous sublayer.
It must be emphasised that the flow here is turbulent, and in this respect
the customary name "laminar sublayer" is unsuitable. The resemblance to
laminar flow lies only in the fact that the mean velocity is distributed accord
ing to the same law as the true velocity would be for laminar flow under the
162 Boundary Layers §42
same conditions. There is, of course, no sharp boundary between the viscous
sublayer and the remainder of the flow, and the concept of the viscous sub
layer is therefore to some extent qualitative.
The longitudinal component v' x of the fluctuating velocity in the viscous
sublayer is of the same order of magnitude as the mean velocity, and in
particular is proportional to y (~ v^yjy ). It therefore follows from the
equation of continuity that the derivative dv'yjdy =  dv' x Jdx is proportional
to y, and so the transverse component v' y of the fluctuating velocity varies as
y 2 ( ~ v *y 2 ly<?)' Next, it follows from the linearity of the equations of motion
in the viscous sublayer (the nonlinear terms being there small compared with
the viscosity terms) that the periods of the turbulent eddies are the same
throughout the thickness of the sublayer. Multiplying these periods by the
fluctuating velocity, we find that the longitudinal distances traversed by the
fluid particles in their fluctuating motion are proportional to y, in order of
magnitude, and the transverse distances are proportional to y*(~ y 2 /yo)
We shall not be further interested in the flow in the viscous sublayer.
Its presence has to be taken into account only in making the appropriate
choice of the constant of integration in (42.4). This constant must be chosen
so that the velocity becomes of the order of v m at distances of the order of yo.
For this to be so, we must take c = log^o, so that u = (vjb) log(y[y ), or
u = (v m fb) log(yv#/ v ). (42.7)
This formula determines (for a certain range of y) the velocity distribution in
the turbulent stream which flows along the surface. This distribution is
called the logarithmic velocity profile.
The argument of the logarithm in formula (42.7) should include a numeri
cal coefficient. However, in the formulae which we shall derive we shall
require only "logarithmic" accuracy. This means that the argument of the
logarithm is supposed large, and we neglect not only terms proportional to
lower powers of the argument but also those involving the logarithm to lower
powers than in the principal term. The introduction of a small numerical
coefficient in the argument of the logarithm in (42.7) is equivalent to adding
a term of the form constant xv m , where the constant is of the order of unity;
this term does not contain the logarithm, and therefore we neglect it. How
ever, it must be borne in mind that the argument of the logarithm in the
formulae derived here is not so large that its logarithm is also very large, and
so the accuracy of the formulae is not very high.
These formulae can be made more exact by introducing a numerical
coefficient in the argument of the logarithm, or, what is the same thing, adding
a constant to the logarithm. These constants, however, cannot be calculated
theoretically, and have to be determined from experimental results. For
example, a more exact formula for the velocity distribution can be written
in the form
u = z; # [240 log0w # /v) + 584]. (42.8)
§43 Turbulent flow in pipes 163
It is not difficult to determine the energy dissipation e per unit mass of
fluid, a is the mean value of the component Ilgy of the momentum flux
density tensor II tt = pvtv k  r)(dv t ldx k + dv k \dxi). Outside the viscous sub
layer, the viscosity term may be omitted, so that a = pv x v y . Introducing
the fluctuating velocity v', we can write v x = u + v' x ; the velocity v y is
itself the fluctuating velocity v' y , since its mean value is zero. The result is
O = pV X Vy = pV' X v'y + pUV'y = pv' X V'y.
Next, the energy flux density in the ydirection is (p + $pv 2 )vy, the viscosity
term being again omitted. Putting in the second term
v* = (u+v'xf+v'tpWf
and averaging, we obtain
pv'y + lp(v' X 2 v'y + V'y* + v' gV y) + p UV'Wy.
Here only the last term need be retained. The reason is that the fluctuating
velocity is of the order of © # , and hence, to logarithmic accuracy, it is small
compared with «. The turbulent fluctuations of the pressure p are of the
order of pv m 2 (cf. (31.4)), and so we can, to the same accuracy, neglect the
corresponding term in the energy flux. Thus we have for the mean energy
flux density puv' x v' y = ua. As we approach the surface, this flux decreases,
because the energy is dissipated. The decrease in the energy flux density
on approaching the surface by a distance d> is <r(dw/dy)d>. This is the amount
of energy converted into heat in a fluid layer of thickness dy and of unit area.
Hence we conclude that the energy dissipation per unit mass is (ojp)Auj6y,
or
e = vflby = {ajpflby. (42.9)
§43. Turbulent flow in pipes
Let us now apply the above results to turbulent flow in a pipe. Near the
walls of the pipe (at distances small compared with its radius a), the surface
may be approximately regarded as plane, and the velocity distribution must
be given by formula (42.7) or (42.8). Since the function logjy varies only
slowly, we can use formula (42.7) to logarithmic accuracy to give the mean
velocity U of the flow in the pipe if we replace y in that formula by a :
U = (v*/b)\og(avM. (43.1)
By U we mean the volume of fluid that passes through a crosssection of the
pipe per unit time, divided by the crosssectional area: U = Qlpna 2 .
In order to relate the velocity U to the pressure gradient Apjl which
maintains the flow (A/> being the pressure difference between the ends of
the pipe, and / its length), we notice that the force on a crosssection of the
164 Boundary Layers §43
flow is 7ra 2 Ap. This force overcomes the friction at the walls. Since the
frictional force per unit area of the wall is a = pv m 2 , the total frictional force
is Tmalpv^. Equating the two forces, we have
Ap/l = 2pv m */a. (43.2)
Equations (43.1) and (43.2) determine, through the parameter v m , the relation
between the velocity of flow in the pipe and the pressure gradient. This
relation is called the resistance law of the pipe. Expressing v m in terms of
Ap// by (43.2), and substituting in (43.1), we obtain the resistance law in the
form
U = V(a&Pl2b*pl) log[(a/vW(a&pl2 P l)]. (43.3)
In this formula it is customary to introduce what is called the resistance
coefficient of the pipe, a dimensionless quantity defined as
laLpll
A = — . (43.4)
The dependence of A on the dimensionless Reynolds number R = 2aU/v is
given in implicit form by the equation
1/VA = 085 log(RVA)055. (43.5)
We have here substituted for b the value (42.2) and added to the logarithm
an empirically determined constant.^ The resistance coefficient determined
by this formula is a slowly decreasing function of the Reynolds number.
For comparison, we give the resistance law for laminar flow in a pipe. Intro
ducing the resistance coefficient in formula (17.10), we obtain
A = 64/R. (43.6)
In laminar flow the resistance coefficient diminishes with increasing Reynolds
number more rapidly than in turbulent flow.
Fig. 22 shows a logarithmic graph of A as a function of R. The steep
straight line corresponds to laminar flow (formula (43.6)), and the less
steep curve (which is almost a straight line also) to turbulent flow. The
transition from the first line to the second occurs, as the Reynolds number
increases, at the point where the flow becomes turbulent; this may occur
for various Reynolds numbers, depending on the actual conditions (the
intensity of the perturbations ; see §29). The resistance coefficient increases
abruptly at the transition point.
t The coefficient of the logarithm in this formula is given to correspond with that in formula
(42.8) for the logarithmic velocity profile. Only in this case does formula (43.5) have the theoretical
significance of being a limiting formula for turbulent flow at sufficiently large values of the Reynolds
number. If the values of the two constants appearing in formula (43.5) are chosen arbitrarily, it can
only be a purely empirical formula for the dependence of A on R. In that case, however, there would
be no reason to prefer it to any other simpler empirical formula which adequately represents the
experimental results.
§43
Turbulent flow in pipes
165
So far we have assumed that the wall surface is fairly smooth. If it is
rough, the formulae obtained above may be somewhat changed. As a measure
of the roughness of the wall, we can take the order of magnitude of the
projections, which we shall denote by d. The relative magnitudes of d
and the thickness yo of the sublayer are of importance. If yo is large compared
with d, the roughness is unimportant; this is what is meant by saying that
the surface is fairly smooth. If yo and d are of the same order of magnitude,
no general formulae can be obtained.
In the opposite limiting case of extreme roughness (d > yo), some general
relations can again be established. In this case we clearly cannot speak of a
viscous sublayer. Turbulent flow occurs around the projections from the
surface, and this flow is characterised by the quantities p,cr,d; the viscosity v,
as usual, cannot appear directly. The velocity of this flow is of the order of
magnitude of v#, the only quantity at our disposal having the dimensions of
velocity. Thus we see that, in flow along a rough surface, the velocity
becomes small (~ v#) at distances y ~ d, instead of v ~ yo as for flow
along a smooth surface. Hence it is clear that the velocity distribution is given
by a formula which is obtained from (42.7) by substituting d for v/v*. Thus
u = (v*lb)log(yld). (43.7)
The formulae for flow in a pipe must be changed similarly. It is sufficient
simply to replace vjv m in them by d. For the resistance law we have, instead of
(43.3), the formula
U = ^{akpjltfpl) log(a/d). (43.8)
The argument of the logarithm is now a constant, and does not involve the
pressure gradient as (43.3) did. We see that the mean velocity is now simply
proportional to the square root of the pressure gradient in the pipe. If we
introduce the resistance coefficient, (43.8) becomes
A = 8&2/log2(a/i) = 14/log2(a/</), (43.9)
i.e. A is a constant and does not depend on the Reynolds number.
166 Boundary Layers §44
§44. The turbulent boundary layer
The fact that we have obtained a logarithmic velocity distribution which
formally holds in all space for planeparallel turbulent flow is due to our
having considered flow along a surface of infinite area. In flow along the
surface of a finite body, only the motion at short distances from the surface —
in the boundary layer — has a logarithmic profile.f We may mention also that
a turbulent boundary layer can exist both under a fluid moving turbulently
in the main stream and under a laminar flow.
The decrease in the mean velocity, both in the turbulent and in the
laminar boundary layer, is due ultimately to the viscosity of the fluid. The
effect of the viscosity appears in the turbulent boundary layer in a rather
unusual manner, however. The manner of variation of the mean velocity in
the layer does not itself depend directly on the viscosity; the viscosity appears
in the expression for the velocity gradient only in the viscous sublayer. The
total thickness of the boundary layer, however, is determined by the viscosity,
and vanishes when the viscosity is zero (see below). If the viscosity were
exactly zero, there would be no boundary layer.
Let us apply the results of §43 to a turbulent boundary layer formed in
flow along a thin flat plate, such as was discussed in §39 with respect to
laminar flow. At the boundary of the turbulent layer, the fluid velocity is
almost equal to the velocity of the main stream, which we denote by U.
To determine this velocity at the boundary we can, however, use formula
(42.7) with logarithmic accuracy, putting the thickness 8 of the boundary
layer instead of y. Equating the two expressions, we obtain
U = (v*/b)log(v*8/v). (44.1)
Here U is a constant parameter for a given flow; the thickness 8, however,
varies along the plate, and v m is therefore also a slowly varying function of x.
Formula (44.1) is inadequate to determine these functions; we need some
other equation, relating v% and 8 to x.
To obtain this, we use the same arguments as in deriving formula (36.3)
for the width of the turbulent wake. As there, the derivative d8/dx must
be of the order of the ratio of the velocity along the jyaxis to that along
the araxis at the boundary of the layer. The latter velocity is of the order
of U, while the former is due to the fluctuating velocity, and is therefore
of the order of v m . Thus dSjdx ~ v%/U, whence
8 ~ v*xjU. (44.2)
t The thickness of the boundary layer increases along the surface of the body in the direction
of flow, according to a law which we shall determine below. This explains why, for flow in a pipe,
the logarithmic profile holds for the whole crosssection of the pipe. The thickness of the boundary
layer at the wall of the pipe increases away from the point of entry of the fluid. At some finite distance
from this point, the boundary layer fills almost the whole crosssection of the pipe. Hence, if we
suppose the pipe sufficiently long and ignore its inlet section, the flow in the whole pipe will be of
the same kind as in the turbulent boundary layer. We may recall that a similar situation occurs for
laminar flow in a pipe. Such a flow obeys Poiseuille's formula for all Reynolds numbers. In Poiseuille
flow the viscosity is important at all distances from the walls, and its effect is never limited to a thin
layer adjoining them.
§44 The turbulent boundary layer 167
Formula (44.1) and (44.2) together determine v m and 8 as functions of the
distance x.f These functions, however, cannot be written explicitly. We
shall express 8 in terms of an auxiliary quantity. Since v # is a slowly varying
function of x, it is seen from (44.2) that the thickness of the layer varies
essentially as x. We may recall that the thickness of the laminar boundary
layer increases as \/x y i.e. more slowly than that of the turbulent boundary
layer.
Let us determine the dependence on x of the frictional force cr acting
on unit area of the plate. This dependence is given by two formulae :
a = P v* 2 , U = {v m jb) Xogip^xjUv).
The latter is obtained by substituting (44.2) in (44.1), and is valid to logarith
mic accuracy. We introduce a drag coefficient c (referred to unit area of the
plate), defined as the dimensionless ratio
c = 2aj P W = 2{vJU) 2 . (44.3)
Then, eliminating v # from the two equations given, we obtain the following
equation, which gives (to logarithmic accuracy) c as an implicit function of x :
V(2#7<0 = log(cRs), R x = Ux/v. (44.4)
To increase the accuracy of this formula, we may add an empirical numerical
constant to the logarithm. Such a formula is,
l/y/c = 17 log(cR x ) + 30. (44.5)
The drag coefficient c given by this formula is a slowly decreasing function
of the distance x.
Finally, let us express the thickness of the boundary layer in terms of the
function c(x). We have v m = \/( a lp) = UVfa) Substituting in (44.2), we
find
8 = constant x x\/c. (44.6)
This formula may be written with the equality sign, of course, only in cases
of a turbulent boundary layer under a laminar flow, when 8 has an exact
significance (the turbulent region being, as always, sharply distinct from
the laminar region). The constant factor in (44.6) has to be determined from
experimental results.
PROBLEMS
Problem 1. Determine from formula (44.5) the total force acting on the two sides of the
plate.
Solution. The required force per unit length of the edge of the plate is
I
= 2 f or dx,
t If there is a laminar boundary layer of considerable extent on the plate, then x must, strictly
speaking, be reckoned as approximately the distance from the point where the laminar layer becomes
turbulent.
168 Boundary Layers §45
where / is the length of the plate. Introducing in place of F the drag coefficient
C = F/ipW.21,
we find
1 r
C =  cdx.
*■ o
If we take only terms containing the logarithm to the highest (first) power, then the above
integral is simply c(Z), the value of c for x = I. In order to obtain a more exact value for C,
corresponding to formula (44.5), we must effect the integration taking account of terms of
the next order, which contain the logarithm to the zero power. To do so, we write
f , I f dc
\ cdx = [#c] — x — dx.
o o "*
The derivative dcfdx is calculated by means of formula (44.5), which we write in the form
c = If A 2 log 2 Bxc, obtaining to the necessary accuracy
2 r 2 1
C = c(l) + = c(l)\l + ,
and so
V V 5 nA \og{BlCje).
Substituting the values of A and B from (44.5), we obtain the following formula, which gives
the total drag coefficient C as a function of the Reynolds number R = Ulfv:
1/VC= l71og(CR)+l3.
For large R, the drag coefficient given by this formula decreases as 1/log 2 R. For the laminar
boundary layer, C decreases as 1/VR (see (39.16)), i.e. more rapidly. Thus we can say that,
for large Reynolds numbers, the frictional force in a turbulent boundary layer is greater than
in a laminar one.
Problem 2. Determine the drag coefficient of a rough plate as a function of the Reynolds
number, for a turbulent boundary layer.
Solution. Substituting in place of the thickness y (^ vjv*) of the laminar sublayer the
dimension d of the projections, we obtain from (44.1) and (44.2) U = (vjb) log(xv JUd).
Introducing the drag coefficient c, we hence have 059J\/c = log(x\/cld). Similarly, the
total drag coefficient for the plate is (again to logarithmic accuracy) 059/t/C = log(l\/Cfd).
We may point out that the drag coefficient for a rough plate is independent of the Reynolds
number.
§45. The drag crisis
From the results obtained in the previous sections we can draw important
conclusions concerning the law of drag for large Reynolds numbers, i.e. the
relation between the drag force acting on the body and the value of R when
the latter is large.
§45 The drag crisis 169
The flow pattern for large R (the only case we shall discuss) has already
been described, and is as follows. Throughout the main body of the fluid
(i.e. everywhere except in the boundary layer, which does not here concern
us) the fluid may be regarded as ideal, with potential flow everywhere
except in the turbulent wake. The width of the wake depends on the position
of the line of separation on the surface of the body. It is important to note
that, although this position is determined by the properties of the boundary
layer, it is found to be independent of the Reynolds number, as we have seen
in §40. Thus we can say that the whole flow pattern for large Reynolds
numbers is almost independent of the viscosity, i.e. of R (so long as the boun
dary layer remains laminar; see below).
Hence it follows that the drag also must be independent of the viscosity.
There remain at our disposal only three quantities: the velocity U of the
main stream, the fluid density p and the dimension / of the body. From these
we can construct only one quantity having the dimensions of force, namely
pTJH 2 . Instead of the squared linear dimension of the body I 2 , we introduce,
as is customarily done, the proportional quantity S, the area of a crosssection
transverse to the direction of flow, putting
F = constant x P U 2 S, (45.1)
where the constant is a number depending only on the shape of the body.
Thus the drag must be (for large R) proportional to the crosssectional area
of the body and to the square of the mainstream velocity. We may recall
for comparison that, for very small R (^ 1), the drag is proportional to the
linear dimension of the body and to the velocity itself (F ~ vplU; see §20).f
It is customary, as we have said, to introduce, in place of the drag force
F, the drag coefficient C defined by C = FJlpU 2 S. This is a dimensionless
quantity, and can depend only on R. Formula (45.1) becomes
C = constant, (45.2)
i.e. the drag coefficient depends only on the shape of the body.
The above behaviour of the drag force cannot continue to arbitrarily
large Reynolds numbers. The reason is that, for sufficiently large R, the
laminar boundary layer (on the surface of the body as far as the line of separa
tion) becomes unstable and hence turbulent. However, the whole boundary
layer does not become turbulent, but only some part of it. The surface of the
body may therefore be divided into three parts : at the front there is a laminar
boundary layer, then a turbulent layer, and finally the region beyond the
line of separation.
The onset of turbulence in the boundary layer has an important effect
on the whole pattern of flow in the main stream. It leads to a considerable
displacement of the line of separation towards the rear of the body (i.e.
t The flow past a bubble of gas is a special case, where the drag remains proportional to U even
for large R; see Problem.
170 Boundary Layers §45
downstream), so that the turbulent wake beyond the body is contracted, as
shown in Fig. 23, where the wake region is shaded.f The contraction of the
turbulent wake leads to a reduction of the drag force. Thus the onset of
turbulence in the boundary layer at large Reynolds numbers is accompanied
by a decrease in the drag coefficient, which falls off by a considerable factor
over a relatively narrow range of Reynolds numbers near 10 5 . We shall call
this phenomenon the drag crisis. The decrease in the drag coefficient is so great
that the drag itself, which for constant C is proportional to the square of the
velocity, actually diminishes with increasing velocity in this range of Reynolds
numbers.
Fig. 23
It may be mentioned that the degree of turbulence in the main stream
affects the drag crisis; the greater the incident turbulence, the sooner the
boundary layer becomes turbulent (i.e. the smaller is R when this happens).
The decrease in the drag coefficient therefore begins at a smaller Reynolds
number, and extends over a wider range of R.
Figs. 24 and 25 give experimentally obtained graphs showing the drag
coefficient as a function of the Reynolds number R = Udjv for a sphere;
Fig. 24 is plotted logarithmically. For very small R ( <^ 1), the drag coefficient
decreases according to C = 24/R (Stokes' formula). The decrease in C
continues more slowly as far as R « 5 x 10 3 , where C reaches a minimum,
beyond which it increases somewhat. In the range of Reynolds numbers
2 x 10 4 to 2 x 10 5 , the law (45.2) holds, i.e. C is almost constant. The drag
crisis occurs for R between 2 x 10 5 and 3 x 10 5 , and the drag coefficient
diminishes by a factor of 4 or 5.
For comparison, we may give an example of flow in which there is no
critical Reynolds number. Let us consider flow past a flat disk in the direction
perpendicular to its plane. In this case the location of the separation is obvious
from purely geometrical considerations: it is clear that separation occurs at
the edge of the disk and does not move from there. Hence, as R increases, the
f For example, in transverse flow past a long cylinder, the onset of turbulence in the boundary
layer moves the point of separation from 95° to 60° (where the azimuthal angle on the cylinder is
measured from the direction of flow).
§45
The drag crisis
171
drag coefficient of the disk remains constant, and there is no drag
crisis.
It must be borne in mind that, for the high velocities at which the
drag crisis occurs, the compressibility of the fluid may begin to have
a noticeable effect. The parameter which characterises the extent of this
effect is the Mack number M = U/c, where c is the velocity of sound; if
M < 1, the fluid may be regarded as incompressible (§10). Since, of the
two numbers M and R, only one contains the dimension of the body, these
two numbers can vary independently.
100
V

60
>
s
\
20
8
C A
2
15
10
0*6
0"3
C
>1
1
2
51
1
D*
1
3 3
1
O 4
r
D b
K
R
Fig. 24
10 5 2i0 5 310 5 410 5 510 5
R
Fig. 25
The experimental data indicate that the compressibility has in general
a stabilising effect on the flow in the laminar boundary layer. When M
172 Boundary Layers §46
increases, the critical value of R increases. For example, when M for a
sphere changes from 03 to 07, the drag crisis is postponed from
R« 4xl0 5 toR« 8x105.
We may also mention that, when M increases, the position of the point
of separation in the laminar boundary layer moves upstream, towards the
front of the body, and this must lead to some increase in the drag.
PROBLEM
Determine the drag force on a gas bubble moving in a liquid at large Reynolds numbers
(V. G. Levich 1949).
Solution. At the boundary between the liquid and the gas the tangential fluid velocity
component does not vanish, but its normal derivative does (we neglect the viscosity of the
gas). Hence the velocity gradient near the boundary will not be particularly high, and there
will be no boundary layer in the sense of §39 ; there will therefore be no separation over almost
the whole surface of the bubble. In calculating the energy dissipation from the volume integral
(16.3) we can therefore use in all space the velocity distribution corresponding to potential
flow past a sphere (§10, Problem 2), neglecting the surface layer of liquid and the very narrow
turbulent wake. Using the formula obtained in §16, Problem, we find
£kin = —q (7— ) 27r# 2 sin0d0 = \2irqRU 2 .
Hence we see that the required dissipative drag isf F = \2tttjRU.
§46. Flow past streamlined bodies
The question may be asked what should be the shape of a body (of a given
crosssectional area, say) for the drag on it resulting from motion in a fluid
to be as small as possible. It is clear from the above that, for this to be so,
the separation must be as far back as possible: the separation must occur
near the rear end of the body, so that the turbulent wake is as narrow as
possible. We know already that the appearance of separation is facilitated by
the presence of a rapid downstream increase in the pressure along the body.
Hence the body must have a shape such that the variation in pressure along
it, where the pressure is increasing, takes place as slowly and smoothly as
possible. This can be achieved by giving the body a shape elongated in the
direction of flow, tapering smoothly to a point downstream, so that the flows
along the two sides of the body meet smoothly without having to go round any
corners or turn through a considerable angle from the direction of the
main stream. At the front end the body must be rounded ; if there were an
angle here, the fluid velocity at its vertex would become infinite (see §10,
Problem 6), and consequently the pressure would increase rapidly down
stream, with separation inevitably resulting.
All these requirements are closely satisfied by shapes of the kind shown
in Fig. 26. The profile shown in Fig. 26b may be, for example, the cross
section of an elongated solid of revolution, or the crosssection of a body with
f The range of applicability of this formula is actually not large, since, when the velocity increases
sufficiently, the bubble ceases to be spherical.
§46 Flow past streamlined bodies 173
a large "span" (we conventionally call such a body a wing). The cross
sectional profile of a wing may be unsymmetrical, as in Fig. 26a. In flow
past a body of this shape, separation occurs only in the immediate neighbour
hood of the pointed end, and consequently the drag coefficient is relatively
small. Such bodies are said to be streamlined.
Fig. 26
The direct friction of the fluid on the surface in the boundary layer is
important in the drag on streamlined bodies. This effect for nonstreamlined
bodies (which were considered in the previous section) is relatively small and
therefore, in practice, of no significance. In the opposite limiting case of
flow parallel to a flat disk, the effect becomes the only source of drag (§39).
In flow past a streamlined wing inclined to the main stream at a small
angle a, called the angle of attack (Fig. 26), a large lift force F y is developed,
while the drag F x remains small, and the ratio F y \F x may therefore reach
large values (~ 10100). This continues, however, only while the angle
of attack is small ( < 10°). For larger angles the drag rises very rapidly, and
the lift decreases. This is explained by the fact that, at large angles of attack,
the body ceases to be streamlined : the point of separation moves a considerable
way towards the front of the body, and the wake consequently becomes
wider. It must be borne in mind that the limiting case of a very thin body, i.e.
a flat plate, is streamlined only for a very small angle of attack; separation
occurs at the leading edge of the plate when it is inclined at even a small
angle to the main stream.
The angle of attack a is, by definition, measured from the position of the
wing for which the lift force is zero. For small angles of attack, we can
expand the lift as a series of powers of a. Taking only the first term, we can
suppose that the force F y is proportional to a. Next, by the same dimensional
arguments as for the drag force, the lift must be proportional to pU 2 . Intro
ducing also the span l z of the wing, we can write
F y = constant x p U 2 cd x l z , (46.1)
where the numerical constant depends only on the shape of the wing and not,
in particular, on the angle of attack. For very long wings, the lift may be
174 Boundary Layers §46
supposed proportional to the span, in which case the constant depends only
on the shape of the crosssection of the wing.
Instead of the lift on the wing, the lift coefficient is often used; it is defined
as
Cy = FylipU%l z . (46.2)
For very long wings, according to what was said above, the lift coefficient is
proportional to the angle of attack, and depends on neither the velocity nor
the span :
Cy = constant x a. (46.3)
To calculate the lift on a streamlined wing by means of Zhukovskii's
formula, it is necessary to determine the velocity circulation Y. This is
done as follows. We have potential flow everywhere outside the wake. In
the present case, the wake is very thin, and occupies on the surface of the
wing only a very small area near its pointed trailing edge. Hence, to determine
the velocity distribution (and therefore the circulation T), we can solve the
problem of potential flow of an ideal fluid round a wing. The existence of the
wake is taken into account by the presence of a tangential discontinuity,
extending into the fluid from the sharp trailing edge of the wing, where the
potential has a discontinuity <j> 2 — <j>\ = Y. As has been shown in §37, the
derivative d<f>fdz also has a discontinuity on this surface, while the derivatives
d<f>jdx and d<f>[dy are continuous. For a wing of finite span, the problem in
this form has a unique solution. The finding of the exact solution is very
complicated, however. The problem has been solved by N. E. KocHiNf
for a wing in the form of a circular disk inclined at a small angle of attack.
If the wing is very long (and has a uniform crosssection), then, regarding
it as infinite in the ^direction, we may regard the flow as twodimensional
(in the xyplane). It is evident from symmetry that the velocity v z = d<f>jdz
along the wing must be zero. In this case, therefore, we must seek a solution
in which only the potential has a discontinuity, its derivatives being con
tinuous; in other words, there is no surface of tangential discontinuity,
and we have simply a manyvalued function <f>(x, y), which receives a finite
increment Y when we go round a closed contour enclosing the profile of the
wing. In this form, however, the problem of twodimensional flow has no
unique solution, since it admits solutions for any given discontinuity of the
potential. To obtain a unique result, we must require the fulfilment of
another condition, first formulated by S. A. Chaplygin in 1909.
This condition, called the ZhukovskiiChaplygin condition, consists in
requiring that the fluid velocity does not become infinite at the sharp trailing
edge of the wing; in this connection we may recall that, when an ideal fluid
flows round an angle, the fluid velocity in general becomes infinite, according
to a power law, at the vertex of the angle (§10, Problem 6). We can say that
t Prikladnaya matematika i mekhamka 4, 3, 1940; 9, 13, 1945.
§47 Induced drag 175
the condition stated implies that the jets coming from the two sides of the
wing must meet smoothly without turning through an angle. When this
condition is fulfilled, of course, the solution of the problem of potential flow
gives a pattern very like the true one, where the velocity is everywhere finite
and separation occurs only at the trailing edge. The solution now becomes
unique and, in particular, the circulation T needed to calculate the lift force
has a definite value.
§47. Induced drag
An important part of the drag on a streamlined wing (of finite span) is
formed by the drag due to the dissipation of energy in the thin turbulent wake.
This is called the induced drag.
It has been shown in §21 how we may calculate the drag force due to the
wake by considering the flow far from the body. Formula (21.1), however,
is not applicable in the present case. According to that formula, the drag is
given by the integral of v x over the crosssection of the wake, i.e. the discharge
through the wake. On account of the thinness of the wake beyond a stream
lined wing, however, the discharge is small in the present case, and may be
neglected in the approximation used below.
As in §21, we write the force F x as the difference between the total fluxes
of the ^component of momentum through the planes x = x\ and x = #2
passing respectively far behind and far in front of the body. Writing the
three velocity components as U+v x , v y , v z , we have for the component Yi xx
of the momentum flux density the expression U xx = p + p(U+v x ) 2 , so that
the drag force is
F * = ( 1/ " 17 )[P + p(U+Vx) 2 ]dydz. (47.1)
x=x t x=x x
On account of the thinness of the wake, we can neglect, in the integral over
the plane x = x±, the integral over the crosssection of the wake, and so
integrate only over the region outside the wake. In that region, however,
we have potential flow, and Bernoulli's equation /> + p(U+v) 2 = p Q + % p U 2
holds, whence
p = popUv x %p(v x 2 +v y z+v z 2 ). (47.2)
Here we cannot neglect the quadratic terms as we did in §21, since it is these
terms which determine the required drag force in the case under considera
tion. Substituting (47.2) in (47.1), we obtain
F * = ( // ~ // )lPo+pU 2 +pUv x +± P (v x Z V«fc 2 )]dyd*.
x=x t x=x t
The difference of the integrals of the constant po + pU 2 is zero; the difference
176 Boundary Layers §47
of the integrals of pUv x is likewise zero, since the mass fluxes
jjpv x dydz
through the front and back planes must be the same (we neglect the discharge
through the wake in the approximation here considered). Next, if we take
the plane x = x% sufficiently far in front of the body, the velocity v on this
plane is very small, so that the integral of \p{v x 2 — v y 2 — v z 2 ) over this plane
may be neglected. Finally, in flow past a streamlined wing, the velocity v x
outside the wake is small compared with v y and v z . Hence we can neglect v x 2
compared with v y 2 + v z 2 in the integral over the plane x = x\. Thus we obtain
Fx = h jj (v y 2 +v z 2 )dydz, (47.3)
where the integration is over a plane x = constant lying at a great distance
behind the body, the crosssection of the wake being excluded from the region
of integration.f
The drag on a streamlined wing calculated in this way can be expressed in
terms of the velocity circulation Y which determines the lift also. To do
this, we first of all notice that, at sufficiently great distances from the body,
the velocity depends only slightly on the coordinate x, and so we can regard
v v{y> %) an( l v z(y, %) as the velocity of a twodimensional flow, supposed
independent of x. It is convenient to use as an auxiliary quantity the stream
function (§10), so that v z = difjjdy, v y = dxjsjdz. Then
>!![{$)'*(&
dydz,
where the integration over the vertical coordinate y is from + oo to ji
and from y% to — oo, where y\ and y^ are the coordinates of the upper and
lower boundaries of the wake (see Fig. 18, §37). Since we have potential flow
(curlv = 0) outside the wake, d 2 ip{dy 2 + d 2 ^jdz 2 = 0. Using the two
dimensional Green's formula, we thus find
F x = lp§<fj{dil>{dn)dl,
where the integral is taken along a contour bounding the region of integration
in the original integral, and djdn denotes differentiation in the direction of the
outward normal to the contour. At infinity ift = 0, and so the integral is taken
t To avoid misunderstanding we should point out the following. Formula (47.3) may give the im
pression that the velocities v y , v z do not decrease in order of magnitude as x increases. This is true
so long as the thickness of the wake is small compared with its width, as we have assumed in deriving
formula (47.3). At very large distances behind the wing, the wake finally becomes so thick that it
becomes approximately circular in crosssection. At this point, formula (47.3) is invalid, and %,
v z diminish rapidly with increasing x.
§47 Induced drag 177
round the crosssection of the wake by the yzptene, giving
Here the integration is over the width of the wake, and the difference in the
brackets is the discontinuity of the derivative 8iff/8y across the wake. Since
8ifsf8y = v z = 8<f)l8z, we have
\8y) 2 Uj/i \8z/ 2 \8zJ1 dz'
so that
F* = ipJ#dr/d*)d*.
Finally, we use a formula from potential theory,
*=J[(S).(S)J— *
where the integration is along a plane contour, r is the distance from dl
to the point where iff is to be found, and the expression in brackets is the
given discontinuity of the derivative of iff in the direction normal to the
contour.) In our case the contour of integration is a segment of the #axis,
so that we can write the value of the function ip(y y z) on the sraxis as
**>4[(E(i),H''"'
= _IW log ,_^<.
2ttJ dz' Si '
Finally, substituting this in F x , we obtain the following formula for the
induced drag:
4tJJ
' 'dl» dr(ar') , , /1J A , , AnA ^
— — — — log \zz'\dzdz' (47 A)
. . dz dz'
(L. Prandtl, 1918). The span of the wing is here denoted by l z = /, and
the origin of z is at one end of the wing.
If all the dimensions in the ^direction are increased by some factor (r
t This formula gives, in twodimensional potential theory, the potential due to a charged plane
contour with a charge density
[(8iffl8n)2(diPldn) 1 ]/2'jT.
178 Boundary Layers §47
remaining constant), the integral (47.4) remains constant, f This shows that
the total induced drag on the wing remains of the same order of magnitude
when its span is increased. In other words, the induced drag per unit length
of the wing decreases with increasing length. $ Unlike the drag, the total
lift force
F y =  P U JFdz (47.5)
increases almost linearly with the span of the wing, and the lift per unit length
is constant.
The following method is convenient for the actual calculation of the inte
grals (47.4) and (47.5). Instead of the coordinate z, we introduce a new
variable 0, defined by
* = /(lcos0) (0 < 6 ^ tt). (47.6)
The distribution of the velocity circulation is written as a Fourier series:
T = 2UI £ A n smnd. (47.7)
The condition that T = at the ends of the wing (z = and /, or 6 =
and it) is then fulfilled.
Substituting the expression (47.7) in (47.5) and effecting the integration
(using the orthogonality of the functions sin and sin n6 for n ^ 1), we
obtain F y = \pU 2 nl 2 A\. Thus the lift force depends only on the first
coefficient in the expansion (47.7). For the lift coefficient (46.2) we have
C y = ttA^i, (47.8)
where we have introduced the ratio A = l\l x of span to width of the wing.
To calculate the drag, we rewrite formula (47.4), integrating once by parts :
p } r , s dlY*') dz'dz
o o
t To avoid misunderstanding, we should mention that it does not matter that the logarithm in
the integrand is increased by a constant when the unit of length is changed. For the integral which
differs from that in (47.4) by having a constant instead of log «— «' is zero, since
j (dr/dz)dz = r,
and the definite integral is zero because T vanishes at the edges of the wake.
% In the limit of infinite span, the induced drag per unit length is zero. In reality, a small amount
of drag remains, determined by the discharge through the wake (i.e. the integral //»* dy dz), which
we have neglected in deriving formula (47.3). This drag includes both the frictional drag and the
remaining part due to dissipation in the wake.
§48 The lift of a thin wing 179
It is easily seen that the integral over z' must be taken as a principal value.
An elementary calculation, with the substitution (47.7),f leads to the following
formula for the induced drag coefficient :
C x = ttX J^nA n 2. (47.10)
The drag coefficient for a wing is defined as
C x = F x l\ P Un x l z , (47.11)
being referred, like the lift coefficient, to unit area in the xsrplane.
PROBLEM
Determine the least value of the induced drag for a given lift and a given span l z — I.
Solution. It is clear from formulae (47.8) and (47.10) that the least value of C x for given
C y (i.e. for given Ay) is obtained if all A n for n ^ 1 are zero. Then
C^min = Cy^jirX. (1)
The distribution of velocity circulation over the span is given by the formula
T= ~Ul x C y ^[z{lz)]. (2)
TTL
If the span is sufficiently large, then the flow round any crosssection of the wing is approxi
mately twodimensional flow round a wing of infinite length and the same crosssection.
In this case we can say that the circulation distribution (2) is obtained for a wing whose shape
in the araplane is an ellipse with semiaxes \l% and \l.
§48. The lift of a thin wing
The problem of calculating the lift force on a wing amounts, by Zhukovskii's
theorem, to that of finding the velocity circulation I\ A general solution of
the latter problem can be given for a thin streamlined wing of infinite span,
the crosssection being the same at every point. { The elegant method of
t In integrating over z' we need the integral
cos«0' 7rsin«0
dd' =
'f  •
J cos 6' — cos 6 sin 8
In integrating over z we use the fact that
sin nd sin md d# = \tt (m = n),
o
= (m ^ n).
J A more detailed account of the theory of twodimensional incompressible flow past a wing is
given by N. E. Kochin, I. A. Kibel* and N. V. Roze, Theoretical Hydromechanics {Teoreticheskaya
gidromekhanika), Part 1, 4th ed., Moscow 1948; L. I. Sedov, Twodimensional Problems of Hydro
dynamics and Aerodynamics (Ploskie zadachi gidrodinamiki i aerodinamiki), Moscow 1950.
180
Boundary Layers
§48
solution given below is due to M. V. Keldysh and L. I. Sedov (1939).
Let y = £i(x) and y = ^(x) be the equations of the lower and upper parts
of the crosssectional profile (Fig. 27). We suppose this profile to be thin, only
slightly curved, and inclined at a small angle of attack to the main stream
(the #axis); that is, both £i, £2 themselves and their derivatives £1', £2'
are small, i.e. the normal to the profile contour is everywhere almost parallel
Fig. 27
to the jaxis. Under these conditions, we may suppose the perturbation v
in the fluid velocity, caused by the presence of the wing, to be everywheref
small compared with the mainstream velocity U. The boundary condition
at the surface of the wing is v y /U = £' for y = £. By virtue of the assump
tions made, we can suppose this condition to hold for y = 0, and not for
y — £. Then we must have on the axis of abscissae between x = and
x == i x == a
v y = Ut,2\x) for y »0 + , v y = U&(x) for y >0. (48.1)
In order to apply the methods of the theory of functions of a complex
variable, we introduce the complex velocity dwfdz = v x — iv y (cf. §10),
which is an analytic function of the variable z = x+iy. In the present case
this function must satisfy the conditions
im(dw/dz) = — U£,2'(x) for y ~+ + ,
im(d«;/d*) =  Z7£i'(«) for y >0, ^^
on the segment (0, a) of the axis of abscissae.
To solve the above problem, we first represent the required velocity
distribution v(x,y) as a sum v = v + + v~ of two distributions having the
following symmetry properties:
©*(*, ~y) = v x {x,y), v y (x, y) = v~ y {x y y),
v + x (x, y) = v+ x (x,y) } v+ y (x, y) = v+ y (x,y).
These properties of the separate distributions v~ and v + do not violate the
equation of continuity or that of potential flow, and, since the problem is
linear, the two distributions may be sought separately.
t Except in a small region near the rounded leading edge of the wing.
§48
The lift of a thin wing
181
(48.4)
The complex velocity is correspondingly represented as a sum
to' = zo'+ + w'_,
and the boundary conditions on the segment (0, a) for the two terms of the
sum are
[imw'+]y^ 0+ = [imw'+] y +o = W(£i+&)>
[im w'_] y ^o + =  [im w'_] y ^o_ = \ U(U'  £ 2 ')
The function zo'_ can be determined at once by Cauchy's formula:
w {*) = ^.ty, d £
Z7rt J <; — z
L
where the integration in the plane of the complex variable £ is along a circle
L of small radius centred at the point £ = z (Fig. 28). The contour L can
Fig. 28
be replaced by a circle C" of infinite radius and a contour C traversed clock
wise ; the latter can be deformed into the segment (0, a) twice over. The
integral along C is zero, since w'(z) vanishes at infinity. The integral
along C gives
to _ =
U r&'(0£i'(0
!
£*
<l.
(48.5)
Here we have used the boundary values (48.4) of the imaginary part of «/_
on the segment (0, a), and the fact that, by the symmetry conditions (48.3),
the real part of «/_ is continuous across this segment.
To find the function «?' + , we have to apply Cauchy's formula, not to this
function itself, but to the product w' + (z)g(z), where g(z) = s/[zj{z— a)]>
and the square root is taken with the plus sign for z — x > a. On the
segment (0, a) of the real axis, the function g(z) is purely imaginary and dis
continuous :g(x+i0) = — g(x— iO) = — i^[xj{a — x)]. It is clear from these
properties of the function g(z) that the imaginary part of the product gio'+
182 Boundary Layers §48
is discontinuous across the segment (0, a), while the real part is continuous,
as with the function w'_. Hence we have, exactly as in the derivation of
formula (48.5),
U f &'(fl + & '(fl
«> +(*)£(*) = z 1 g(£+i0)dl
Lit J S~Z
Collecting the above expressions, we have the following formula for the
velocity distribution in flow past a thin wing:
:d£
^ = _ _^_ /fZf f &'(fl +&'(!) /_jr_
d# 2irt'V # J £# N a£
o
"■f*'<fl*'®«. (48.6)
Near the rounded leading edge (i.e. for z > 0), this expression in general
becomes infinite, the approximation used above being invalid in this region.
Near the pointed trailing edge (i.e. for z > a), the first term in (48.6) is
finite, but the second term becomes infinite, though only logarithmically, f
This logarithmic singularity is due to the approximation used, and is removed
by a more exact treatment; there is no powerlaw divergence at the trailing
edge, in accordance with the ZhukovskiiChaplygin condition. The fulfilment
of this condition is achieved by an appropriate choice of the function g(z)
used above.
Formula (48.6) immediately enables us to determine the velocity circulation
T round the wing profile. According to the general rule (see §10), T is
given by the residue of the function w'{z) at its simple pole z = 0. The
required residue is easily found as the coefficient of \jz in an expansion of
to\z) in powers of \\z about the point at infinity: dzojdz = rj2Triz+... ,
and r is given by the simple formula
a
r= uj(K+&)J^&. (48.7)
o
We may point out that only the sum of the functions £i and £2 appears here.
The lift force is unchanged if the thin wing is replaced by a bent plate whose
shape is given by the function (£i+ £2).
For example, for a wing in the form of a thin plate of infinite length,
inclined at a small angle of attack a, we have £1 = £2 = a(a — x), and for
mula (48.7) gives F — —ircnaU. The lift coefficient for such a wing is
C y = pXJTjyWa = 2™.
t This divergence disappears if £ x and £ 2 vanish as (a — x)k, k > 1, near the trailing edge, i.e. if the
point at the trailing edge is a cusp.
CHAPTER V
THERMAL CONDUCTION IN FLUIDS
§49. The general equation of heat transfer
It has been mentioned at the end of §2 that a complete system of equations
of fluid dynamics must contain five equations. For a fluid in which processes
of thermal conduction and internal friction occur, one of these equations is,
as before, the equation of continuity, and Euler's equations are replaced by
the NavierStokes equations. The fifth equation for an ideal fluid is the
equation of conservation of entropy (2.6). In a viscous fluid this equation
does not hold, of course, since irreversible processes of energy dissipation
occur in it.
In an ideal fluid the law of conservation of energy is expressed by
equation (6.1):
a
— ($pv 2 + pe) = diy\fiv(^v 2 + w)].
ot
The expression on the left is the rate of change of the energy in unit volume of
the fluid, while that on the right is the divergence of the energy flux density.
In a viscous fluid the law of conservation of energy still holds, of course:
the change per unit time in the total energy of the fluid in any volume must
still be equal to the total flux of energy through the surface bounding that
volume. The energy flux density, however, now has a different form.
Besides the flux p\(^v 2 + zo) due to the simple transfer of mass by the motion
of the fluid, there is also a flux due to processes of internal friction. This
latter flux is given by the vector v»o', with components (;<</{* (see §16).
There is, moreover, another term that must be included in the energy flux.
If the temperature of the fluid is not constant throughout its volume, there
will be, besides the two means of energy transfer indicated above, a transfer of
heat by what is called thermal conduction. This signifies the direct molecular
transfer of energy from points where the temperature is high to those where
it is low. It does not involve macroscopic motion, and occurs even in a fluid
at rest.
We denote by q the heat flux density due to thermal conduction. The
flux q is related to the variation of temperature through the fluid. This
relation can be written down at once in cases where the temperature gradient
in the fluid is not large; in phenomena of thermal conduction we are almost
always concerned with such cases. We can then expand q as a series of powers
of the temperature gradient, taking only the first terms of the expansion. The
constant term is evidently zero, since q must vanish when grad T does so.
Thus we have
q = /cgradT. (49.1)
183
184 Thermal Conduction in Fluids §49
The constant k is called the thermal conductivity. It is always positive, as
we see at once from the fact that the energy flux must be from points at a
high temperature to those at a low temperature, i.e. q and grad T must be
in opposite directions. The coefficient k is in general a function of tempera
ture and pressure.
Thus the total energy flux in a fluid when there is viscosity and thermal
conduction is p\{\v 2 + w)\»a'  k grad T. Accordingly, the general law
of conservation of energy is given by the equation
d
— (%pv 2 + pe) =»  div[pv(%v 2 + w)  v • a'  k grad T] . (49.2)
This equation could be taken to complete the system of fluidmechanical
equations of a viscous fluid. It is convenient, however, to put it in another
form by transforming it with the aid of the equations of motion. To do so, we
calculate the time derivative of the energy in unit volume of fluid, starting
from the equations of motion. We have
d dp by de dp
Substituting for dpjdt from the equation of continuity and for dw\dt from
the NavierStokes equation, we have
d
— (p^ 2 +pe) = — $v 2 di\(pv) — pv»grad^ 2 — v»grad/> +
dt
+ vi — + p—  e div(pv).
cxjc at
Using now the thermodynamic relation de = Tds— p dV = Tds+(plp 2 )dp,
we find
de ds P dp ds p
— =T— + — —=T — div(pv).
dt dt p 2 dt et p 2 v '
Substituting this and introducing the heat function w = e+p[p, we obtain
d
— (Ipv 2 + pe) — — (%v 2 + w) div(pv) — pv • grad^w 2 — v • grad/) +
dt
ds da ilc
+ pT Wi .
dt dxjc
Next, from the thermodynamic relation dw = Tds+dp/p we have
grad/) = p grad tv — p T grad s. The last term on the right of the above
equation can be written
da iv d dvi di)*
= —— (vio'iie)  cr'ac — = div(va')ff'i
dxic dxjc dxjc dxjc
§49 The general equation of heat transfer 1 85
Substituting these expressions, and adding and subtracting div(/c grad T),
we obtain
d
— (%pv 2 +pe) == — div[pv(%v 2 + w)— v»o' — KgradT] +
dt
I ds \ dvi
+ P T[ — + vgrad*  a' ik —   div(ic grad T). (49.3)
\ dt / dxjc
Comparing this expression for the time derivative of the energy in unit
volume with (49.2), we have
/ ds \ dvi
pTI — + vgrads = a' ilc —  + div(* grad T). (49.4)
\ 8t I dxjc
This equation is called the general equation of heat transfer. If there is no
viscosity or thermal conduction, the righthand side is zero, and the equation
of conservation of entropy (2.6) for an ideal fluid is obtained.
The following interpretation of equation (49.4) should be noticed. The
expression on the left is just the total time derivative ds/dt of the entropy,
multiplied by pT. The quantity ds/dt gives the rate of change of the entropy
of a unit mass of fluid as it moves about in space, and T ds/dt is therefore
the quantity of heat gained by this unit mass in unit time, so that pT dsjdt
is the quantity of heat gained per unit volume. We see from (49.4) that
the amount of heat gained by unit volume of the fluid is therefore
v'ik dvildxic + div(K grad T).
The first term here is the energy dissipated into heat by viscosity, and the
second is the heat conducted into the volume concerned.
We expand the term a'acdvijdxk in (49.4) by substituting the expression
(15.3) for o' ik . We have
dvi dvi / dvi dvjc dvi \ dvi dvi
a'ik = f] ( 1 I §ik I + LfT^ik— — •
dxjc dxjc \dxjc 8xi oxi 1 dxjc oxi
It is easy to verify that the first term may be written as
dvi dvjc ^ dvi\'
and the second is
\ dxjc dxt dxi /
dvi dvi dvi dvi
OXlc OXl OXi OX\
Thus equation (49.4) becomes
ds \ I dvi dvjc dvi \ 2
 + vgrads = div( K grad T)+± v —  + —   fS i& — +
at I \ dxjc oxi oxi I
+ £(divv) 2 . (49.5)
186 Thermal Conduction in Fluids §49
The entropy of the fluid increases as a result of the irreversible processes
of thermal conduction and internal friction. Here, of course, we mean not the
entropy of each volume element of fluid separately, but the total entropy of the
whole fluid, equal to the integral
J* psdV.
The change in entropy per unit time is given by the derivative
d[ j P sdV]Jdt = j [8{ps)/dt]dV.
Using the equation of continuity and equation (49.5) we have
8( P s) 8s 8 P 1
+ s— = s div(pv)pvgrads H div(/c grad T) +
8t 8t 8t ° T
7) I 8v t 8v k 8vi\* I
The first two terms on the right together give div(psv). The volume
integral of this is transformed into the integral of the entropy flux psv
over the surface. If we consider an unbounded volume of fluid at rest at
infinity, the bounding surface can be removed to infinity; the integrand in the
surface integral is then zero, and so is the integral itself. The integral of
the third term on the right is transformed as follows:
Assuming that the fluid temperature tends sufficiently rapidly to a constant
value at infinity, we can transform the first integral into one over an infinitely
remote surface, on which grad T = and the integral therefore vanishes.
The result is
d f C K(gradT) 2 r r> / 8vi 8v k 8vi\ 2
+ f— (divv)2dF. (49.6)
The first term on the right is the rate of increase of entropy owing to thermal
conduction, and the other two terms give the rate of increase due to internal
friction. The entropy can only increase, i.e. the sum on the right of (49.6)
must be positive. In each term, the integrand may be nonzero even if the
other two integrals vanish. Hence it follows that the second viscosity
coefficient £ is positive, as well as k and 77, which we already know are positive.
It has been tacitly assumed in the derivation of formula (49.1) that the
heat flux depends only on the temperature gradient, and not on the pressure
gradient. This assumption, which is not evident a priori, can now be justified
§49 The general equation of heat transfer 187
as follows. If q contained a term proportional to grad/>, the expression
(49.6) for the rate of change of entropy would include another term having
the product grad/>*grad T in the integrand. Since the latter might be either
positive or negative, the time derivative of the entropy would not necessarily
be positive, which is impossible.
Finally, the above arguments must also be refined in the following respect.
Strictly speaking, in a system which is not in thermodynamic equilibrium,
such as a fluid with velocity and temperature gradients, the usual definitions
of thermodynamic quantities are no longer meaningful, and must be modified.
The necessary definitions are, firstly, that p, e and v are defined as before:
p and pe are the mass and internal energy per unit volume, and v is the
momentum of unit mass of fluid. The remaining thermodynamic quantities
are then defined as being the same functions of p and e as they are in
thermal equilibrium. The entropy s = s(p, e), however, is no longer the true
thermodynamic entropy: the integral
/'
sdV
will not, strictly speaking, be a quantity that must increase with time. Never
theless, it is easy to see that, for small velocity and temperature gradients, $
is the same as the true entropy in the approximation here used. For, if there
are gradients present, they in general lead to additional terms (besides
s(p, e)) in the entropy. The results given above, however, can be altered only
by terms linear in the gradients (for instance, a term proportional to the scalar
div v). Such terms would necessarily take both positive and negative values.
But they ought to be negative definite, since the equilibrium value $ = s(p, e)
is the maximum possible value. Hence the expansion of the entropy in powers
of the small gradients can contain (apart from the zeroorder term) only
terms of the second and higher orders.
Similar remarks should have been made in §15 (cf. the first footnote to that
section), since the presence of even a velocity gradient implies the absence of
thermodynamic equilibrium. The pressure p which appears in the expression
for the momentum flux density tensor in a viscous fluid must be taken to be
the same function p = p(p, e) as in thermal equilibrium. In this case p
will not, strictly speaking, be the pressure in the usual sense, viz. the normal
force on a surface element. Unlike what happens for the entropy (see
above), there is here a resulting difference of the first order with respect to
the small gradient; we have seen that the normal component of the force
includes, besides p, a term proportional to div v (in an incompressible fluid,
this term is zero, and the difference is then of higher order).
Thus the three coefficients 17, £, k which appear in the equations of
motion of a viscous conducting fluid completely determine the fluidmechani
cal properties of the fluid in the approximation considered (i.e. when the
higherorder space derivatives of velocity, temperature, etc. are neglected).
The introduction of any further terms (for example, the inclusion in the mass
flux density of terms proportional to the gradient of density or temperature)
188 Thermal Conduction in Fluids §50
has no physical meaning, and would mean at least a change in the definition
of the basic quantities; in particular, the velocity would no longer be the
momentum of unit mass of fluid. f
§50. Thermal conduction in an incompressible fluid
The general equation of thermal conduction (49.4) or (49.5) can be con
siderably simplified in certain cases. If the fluid velocity is small compared
with the velocity of sound, the pressure variations occurring as a result of
the motion are so small that the variation in the density (and in the other
thermodynamic quantities) caused by them may be neglected. However, a
nonuniformly heated fluid is still not completely incompressible in the
sense used previously. The reason is that the density varies with the tem
perature; this variation cannot in general be neglected, and therefore, even
at small velocities, the density of a nonuniformly heated fluid cannot be
supposed constant. In determining the derivatives of thermodynamic
quantities in this case, it is therefore necessary to suppose the pressure con
stant, and not the density. Thus we have
8s I 8s \ 8T I 8s \
It = \8f) v ~8~t y gra Sz= l"ar/ p gra '
and, since T(ds/8T) p is the specific heat at constant pressure c py we obtain
T8sj8t = c p 8T/8t, Tgrads = c p grad T. Equation (49.4) becomes
/ 8T \ 8vi
pc p \ + vgrad T = div(« grad T) + a' ik . (50.1)
\ 8t / 8xjc
If the density is to be supposed constant in the equations of motion for
a nonuniformly heated fluid, it is necessary that the fluid velocity should be
small compared with that of sound, and also that the temperature differences
in the fluid should be small. We emphasise that we mean the actual values of
the temperature differences, not the temperature gradient. The fluid may
then be supposed incompressible in the usual sense ; in particular, the equation
of continuity is simply div v = 0. Supposing the temperature differences
small, we neglect also the temperature variation of r), k and c p , supposing
them constant. Writing the term o'yc dvt/dxjc as in (49.5), we obtain the
t Worse still, the inclusion of such terms may violate the necessary conservation laws. It must
be borne in mind that, whatever the definitions used, the mass flux density j must always be the
momentum of unit volume of fluid. For j is denned by the equation of continuity,
d P /8t + div} = 0;
multiplying this by r and integrating over the fluid volume, we have
d(j P rdV)/dt = jjdV,
and since the integral Jpr dV determines the position of the centre of mass, it is clear that the integral
J" j dV is the momentum.
§50 Thermal conduction in an incompressible fluid 189
equation of heat transfer in an incompressible fluid in the following com
paratively simple form:
8T v I dvi dvir\ 2
— + v.gradr = x Ar+— — + — ), (50.2)
where v — f\\p is the kinematic viscosity, and we have written k in terms of
the thermometric conductivity, defined as
x = Klpc p . (50.3)
The equation of heat transfer is particularly simple for an incompressible
fluid at rest, in which the transfer of energy takes place entirely by thermal
conduction. Omitting the terms in (50.2) which involve the velocity, we have
simply
8T/dt = x^ T  (50.4)
This equation is called in mathematical physics the equation of thermal
conduction or Fourier's equation. It can, of course, be obtained much more
simply without using the general equation of heat transfer in a moving
fluid. According to the law of conservation of energy, the amount of heat
absorbed in some volume in unit time must equal the total heat flux into this
volume through the surface surrounding it. As we know, such a law of
conservation can be expressed as an "equation of continuity" for the amount of
heat. This equation is obtained by equating the amount of heat absorbed in
unit volume in unit time to minus the divergence of the heat flux density.
The former is pc p dTjdt; we must take the specific heat c p , since the pressure
is of course constant throughout a fluid at rest. Equating this to — div q
= kAT, we have equation (50.4).
It must be mentioned that the applicability of the thermal conduction
equation (50.4) to fluids is actually very limited. The reason is that, in
fluids in a gravitational field, even a small temperature gradient usually
results in considerable motion (convection; see §56). Hence we can actually
have a fluid at rest with a nonuniform temperature distribution only if the
direction of the temperature gradient is opposite to that of the gravitational
force, or if the fluid is very viscous. Nevertheless, a study of the equation
of thermal conduction in the form (50.4) is very important, since processes of
thermal conduction in solids are described by an equation of the same form.
We shall therefore consider it in more detail in §§51 and 52.
If the temperature distribution in a nonuniformly heated medium at rest
is maintained constant in time (by means of some external source of heat),
the equation of thermal conduction becomes
AT=0. (50.5)
Thus a steady temperature distribution in a medium at rest satisfies Laplace's
equation. In the more general case where k cannot be regarded a constant,
we have in place of (50.5) the equation
div(/c grad T) = 0. (50.6)
190 Thermal Conduction in Fluids §50
If the fluid contains external sources of heat (for example, heating by
an electric current), the equation of thermal conduction must correspondingly
contain another term. Let Q be the quantity of heat generated by these
sources in unit volume of the fluid per unit time; Q is, in general, a function
of the coordinates and of the time. Then the heat balance equation, i.e.
the equation of thermal conduction, is
P c p dT/dt= kAT+Q. (50.7)
Let us write down the boundary conditions on the equation of thermal
conduction which hold at the boundary between two media. First of all, the
temperatures of the two media must be equal at the boundary:
7i = T 2 . (50.8)
Furthermore, the heat flux out of one medium must equal the heat flux
into the other medium. Taking a coordinate system in which the part of the
boundary considered is at rest, we can write this condition as/ci grad 2V df
= k 2 grad T 2 • df for each surface element df. Putting grad T» df = (dTjdrfdf,
where BTjdn is the derivative of T along the normal to the surface, we obtain
the boundary condition in the form
/ci 07i/a» = K 2 8T 2 l8n. (50.9)
If there are on the surface of separation external sources of heat which
generate an amount of heat £) (s) on unit area in unit time, then (50.9) must
be replaced by
K 1 dT 1 l8nK 2 dT 2 ldn = £)<«>. (50.10)
In physical problems concerning the distribution of temperature in the
presence of heat sources, the strength of the latter is usually given as a
function of temperature. If the function Q{T) increases sufficiently rapidly
with T, it may be impossible to establish a steady temperature distribution
in a body whose boundaries are maintained in fixed conditions (e.g. at a given
temperature). The loss of heat through the outer surface of the body is
proportional to some mean value of the temperature difference T— To between
the body and the external medium, regardless of the law of heat generation
within the body; it is clear that, if the generation of heat increases sufficiently
rapidly with temperature, the loss of heat may be inadequate to achieve an
equilibrium state.
The impossibility of establishing a steady thermal state forms the basis of
the thermal theory of explosions developed by N. N. Semenov (1928): if the
rate of an exothermic combustion reaction increases sufficiently rapidly with
temperature, the impossibility of a steady distribution leads to a rapid
nonsteady ignition of the substance and an acceleration of the reaction into a
thermal explosion. A quantitative theory, for the case where the heat
§50 Thermal conduction in an incompressible fluid 191
generation is an exponential function of temperature, has been given by
D. A. FrankKamenetski! (see Problem l).f
PROBLEMS
Problem 1. Heat sources of strength Q = Q tt e cciT  T J per unit volume are distributed in a
layer of material bounded by two parallel infinite planes, which are kept at a constant tem
perature T . Find the condition for a steady temperature distribution to be possible.
Solution. The equation for steady heat conduction is here
K&Tldx* = £ oC «<ZMr.),
with the boundary conditions T = T for x = and x = 21 (21 being the thickness of the
layer). We introduce the dimensionless variables r = ct(T—T ) and £ = xjl. Then
t" + A<* = 0, A = £ aZ 2 //o
Integrating this equation once (after multiplying by 2t'), we find
t'2 = 2\{e^e%
where t is a constant, which is evidently the maximum value of t; by symmetry, this value
must be attained halfway through the layer, i.e. for £ — 1. Hence a second integration, with
the condition t = for £ = 0, gives
1 r &r r
V(2A) J V^er) = J df==1 '
Effecting the integration, we have
e~* T . coshM r . = \Z(i*). (1)
The function A(t ) determined by this equation has a maximum A = A cr for a definite value
T o — T o.or>" if A > A CT , there is no solution satisfying the boundary conditions.^ The numerical
values are A„ = 088, T 0>cr = 1 '2.ft
Problem 2. A sphere is immersed in a fluid at rest, in which a constant temperature
gradient is maintained. Determine the resulting steady temperature distribution in the fluid
and the sphere.
Solution. The temperature distribution satisfies the equation A T = in all space, with
the boundary conditions
7i = T 2 , /ci dTxjdr = k 2 dT 2 /dr
for r = R (where R is the radius of the sphere ; quantities with the suffixes 1 and 2 refer to
the sphere and the fluid respectively), and grad T = A at infinity, where A is the given
f The rate of explosive combustion reactions, and therefore the rate of heat generation, depend
on temperature roughly as e  v l BT , the constant U being large. Frank KamenetskiI has shown that,
to investigate the conditions for a thermal explosion to occur, we must consider the course of the
reaction when the ignition of the substance is comparatively slow, and therefore replace e~ u l BT by
euvir, e v(TT tt )iRT a * y w here T is the external temperature. A more detailed discussion is given in the
book by D. A. Frank KamenetskiI, Diffusion and Heat Exchange in Chemical Kinetics Princeton
1955.
% Only the smaller of the two roots of equation (1) for A < Acr corresponds to a stable temperature
distribution.
ff The corresponding values for a spherical region (of radius Z) are Acr — 332, T 0>C r — 147, and for
an infinite cylinder A cr = 2 00, T 0lC r = 136.
192 Thermal Conduction in Fluids §51
temperature gradient. By the symmetry of the problem, A is the only vector which can
determine the required solution. Such solutions of Laplace's equation are constant X At
and constant X A • grad(l /r). Noticing also that the solution must remain finite at the centre
of the sphere, we seek the temperatures T x and T 2 in the forms
7i = Cl A.r, T 2 = c 2 A.r/r3+A.r.
The constants c t and c 2 are determined from the conditions for r = R, the result being
3/C2 I" K2 — K1 / R \ 3 1
Ti =
' »['♦££(')>«
K± + 2/C2
§51. Thermal conduction in an infinite medium
Let us consider thermal conduction in an infinite medium at rest. The
most general problem of this kind is as follows. The temperature distribution
is given in all space at the initial instant t = :
T = To(x,y,z) for t = 0,
where To is a given function of the coordinates. It is required to determine
the temperature distribution at all subsequent instants.
We expand the required function T as a Fourier integral with respect to
the coordinates:
T = j 7k(*)exp(*k.r)d3k, d3k = <\k x dk y dk z , (51.1)
where the expansion coefficients are given by
n{t) = (2tt)3 J r(^/,s',*) ex P(^"rW', dV = dx'&y'&z'.
Substituting the expression (51.1) in equation (50.4), we obtain
dTk
f l^ + & 2 xrjexp(*k.r)d3k = 0,
whence
dTJdt+kz x T k = 0.
This equation gives T k as a function of time :
T k = exp(— k 2 xt)T 0k .
Substituting this in (51.1), we find
T = J" r 0k exp(#y) exp(/k.r)d3k. (51.2)
Since we must have T = To(x, y, z) for / = 0, it is clear that the
Tok are the expansion coefficients of the function To(x, y, z) as a Fourier
integral:
T 0k = (2tt)3 J r (^/,*')exp(tk.r')dF.
§51 Thermal conduction in an infinite medium 193
Finally, substituting this in (51.2), we obtain
T = (2tt)3 J J T (x',y',z') exp(^) exp\ik.(rr')]dV &k.
The integral over k is the product of three simple integrals, each of the form
00 00
J exp( — k x 2 xt) exp [ik x (x — x')] dk x = J exp( — k x 2 xt) cos k x (x —x') dk x ;
—00 —00
the similar integral with sin in place of cos is zero, since the sine function
is odd. Using the formula
00
I exp( — ax 2 ) cos /tod*; = ■\/( 7T / a ) ex P( — /? 2 /4a) (a > 0),
—oo
we have finally
T(x,y,z,t) = — —  JV (*', /, *') x
exp{ [{x  x'f + (y y'f + (z z'f]l4 x t} dV. (51.3)
This formula gives the complete solution of the problem ; it determines the
temperature distribution at any instant in terms of the given initial distri
bution.
If the initial temperature distribution is a function of only one coordinate,
x, then we can integrate over y' and z' in (51.3) and obtain
T(x, t) = —— T 7o(*') exp[(**') 2 /4x'] d*'. (51.4)
VW) oo
At time t = 0, let the temperature be zero in all space except for an infinitely
thin layer at the plane x = 0, where it is infinite in such a way that the total
quantity of heat (proportional to $To(x)dx) is finite. Such a distribution can
be represented by a delta function: To(x) = constant x8(x). The integration
in formula (51.4) then amounts to replacing x' by zero, the result of which is
1
T(x, t) = constant x exp(  # 2 /4x*) (51.5)
Similarly, if at the initial instant a finite quantity of heat is concentrated
at a point (the origin), the temperature distribution at subsequent instants
is given by the formula
1
T(r, t) = constant x — — exp(  r 2 /4 x t), (51.6)
S{7Txty
where r is the distance from the origin. In the course of time, the temperature
194
Thermal Conduction in Fluids
§51
at the point r = decreases as H. The temperature in the surrounding
space rises correspondingly, and the region where the temperature is appre
ciably different from zero expands (Fig. 29). The manner of this expansion
is determined principally by the exponential factor in (51.6). We see that
the order of magnitude / of the dimension of this region is given by I 2 fat ~ 1,
whence
i ~ V(xt),
i.e. / increases as the square root of the time.
(51.7)
Formula (51.7) can also be interpreted in a somewhat different way. Let /
be the order of magnitude of the dimension of a body. Then we can say that,
if the body is heated nonuniformly, the order of magnitude r of the time
required for the temperature to become more or less the same throughout
the body is
r ~ J 2 /*. (51.8)
The time t, which may be called the relaxation time for thermal conduction,
is proportional to the square of the dimension of the body, and inversely
proportional to the thermometric conductivity.
The thermal conduction process described by the formulae obtained above
has the property that the effect of any perturbation is propagated instan
taneously through all space. It is seen from formula (51.5) that the heat
§51 Thermal conduction in an infinite medium 195
from a point source is propagated in such a manner that, even at the next
instant, the temperature of the medium is zero only at infinity. This property
holds also for a medium in which the thermometric conductivity x depends on
the temperature, provided that x does not vanish anywhere. If, however, x
is a function of temperature which vanishes when T = 0, the propagation of
heat is retarded, and at each instant the effect of a given perturbation extends
only to a finite region of space (we suppose that the temperature outside this
region can be taken as zero). This result, as well as the solution of the
following Problems, is due to Ya. B. Zel'dovich and A. S. Kompaneets
(1950).
PROBLEMS
Problem 1. The specific heat and thermal conductivity of a medium vary as powers of the
temperature, while its density is constant. Determine the manner in which the tempera
ture tends to zero near the boundary of the region which, at a given instant, has received
heat propagated from an arbitrary source (the temperature outside that region being zero).
Solution. If k and c P vary as powers of the temperature, the same is true of the thermo
metric conductivity x and of the heat function
ZO
= j CpdT
(we omit a constant in w). Hence we can put x — aW n , where we denote by W = pw the
heat function per unit volume. Then the thermal conduction equation
pc p BTjdt = div(*c grad T)
becomes
dW/dt = a di\(Wn grad W). (1)
During a short interval of time, a small portion of the boundary of the region may be
regarded as plane, and its rate of displacement in space, v, may be supposed constant. Accord
ingly, we seek a solution of equation (1) in the form W = W(x —vt), where x is the coordinate
in the direction perpendicular to the boundary. We have
vdWjdx = ad(W« dW/dx)ldx, (2)
whence we find, after two integrations, that W vanishes as
W ~ *i/» (3)
where \x\ is the distance from the boundary of the heated region. This also confirms our
conclusion that, if n > 0, the heated region has a boundary outside which W and T are
zero. If n < 0, then equation (2) has no solution vanishing at a finite distance, i.e. the heat
is distributed through all space at every instant.
Problem 2. A medium like that described in Problem 1 has, at the initial instant, an amount
of heat Q per unit area concentrated on the plane * = 0, while T = everywhere else. Deter
mine the temperature distribution at subsequent instants.
Solution. In the onedimensional case, equation (1) of Problem 1 is
dw a / 8W\
= a —\Wn . (1)
dt dx\ 8x ]
From the parameters Q and a and variables * and t at our disposal, we can form only one
dimensionless combination,
£ = x/(Qnat)W+M; (2)
196 Thermal Conduction in Fluids §52
Q and a have the dimensions erg/cm 8 and (cm a /sec)(cm 3 /erg)». Hence the required function
W(x, t) must be of the form
^=(2 2 M 1/(2+ »m (3)
where the dimensionless function /(0 is multiplied by a quantity having the dimensions
erg/cm 3 . With this substitution, equation (1) gives
d / d/\ dY
This ordinary differential equation has a simple solution which satisfies the conditions of the
problem, namely
f(i) = [M& 2  2 )/(2 + n)]i/», ( 4)
where £ is a constant of integration.
For n > 0, this formula gives the temperature distribution in the region between the planes
x = ±* corresponding to the equation £ = ±$ ; outside this region, W = 0. Hence it
follows that the heated region expands with time in a manner given by x = constant X i x /( 2 +»).
The constant £ is determined by the condition that the total amount of heat is constant:
Xo £o
Q j Wdx = Q J7(fld£ (5)
— #0 ~~ So
whence we have
= (2+^21* r»(j+i/*»)
n7T n/2 r»(l/w) ' w
For n — — v < 0, we write the solution in the form
Here the heat is distributed through all space, and at large distances W decreases as x~ ilv .
This solution is valid only for v < 2 ; for v > 2, the normalisation integral (5) (which now
extends to ± oo) diverges, which means physically that the heat is conducted instantaneously
to infinity. For v < 2, the constant £ in (7) is given by
2(2„k/2 r(i/ y j)
h = — ; — j^m W
Finally, f or n > we have £ » 2/V«, and the solution given by formula (3) of Problem 1
(1), and (4) is '
( Q l x 2 \ 1/n \ o
w  feferw 1 = 2^br*«
in agreement with formula (51.5).
§52. Thermal conduction in a finite medium
In problems of thermal conduction in a finite medium, the initial tempera
ture distribution does not suffice to determine a unique solution, and the
boundary conditions at the surface of the medium must also be given.
§52 Thermal conduction in a finite medium 197
Let us consider thermal conduction in a halfspace (x > 0), beginning
with the case where a given constant temperature is maintained on the
bounding plane x = 0. We may arbitrarily take this temperature as zero.
At the initial instant, the temperature distribution throughout the medium is
given, as before. The boundary and initial conditions are therefore
T = for * = 0; T = T (x,y,z) for t = and * > 0. (52.1)
The solution of the thermal conduction equation with these conditions can,
by means of the following device, be reduced to the solution for a medium
infinite in all directions. We imagine the medium to extend on both sides of
the plane x = 0, the temperature distribution for t = and x < being
given by — To. That is, the temperature distribution at the initial instant
is given in all space by an odd function of x :
?o(  x, y, z)=  T (x,y, z). (52.2)
It follows from equation (52.2) that 7o(0, y, z) = — 7o(0, y, z) = 0, i.e. the
necessary boundary condition (52.1) is automatically satisfied for t = 0,
and it is evident from symmetry that it will continue to be satisfied for all t.
Thus the problem is reduced to the solution of equation (50.4) in an
infinite medium with an initial function To(x,y, z) which satisfies (52.2),
and without boundary conditions. Hence we can use the general formula
(51.3),. We divide the range of integration over x' in (51.3) into two parts,
from — oo to and from to oo. Using the relation (52.2), we then have
OO OO 00
r(w<) = iiiJJI r ° (W):
—oo—oo
{exp[(x—x'f/4xt]exp[(x+x') 2 l4xt]} x
exp {  [(y  y'f + (z z'f]l4 x t] oV Ay' dz'. (52.3)
This formula gives the solution of the problem, since it determines the tem
perature throughout the medium, i.e. for all x > 0.
If the initial temperature distribution is a function of x only, formula
(52.3) becomes
1 °°
T (*'') = ^TTT^ \ 71 o(^){exp[(^^)2 / 4^ ] _e X p[_^ + ^)2/4^]}d^.
(52.4)
As an example, let us consider the case where the initial temperature is a
given constant everywhere except at x = 0. Without loss of generality, this
constant may be taken as — 1. The temperature on the plane x = is always
zero. The appropriate solution is obtained at once by substituting
198
Thermal Conduction in Fluids
§52
Tq{x) = — 1 in (52.4). The integral in (52.4) is the sum of two integrals, in
each of which we change the variables as in £ = (x' — x)/2^(xt). We then
obtain for T(x, t) the expression
T{x,t) = Kerf[*/2V(*0]erf[*/2V(xO]}>
where the function erf x is defined as
2 r ,
erf# = £~^d£,
\fn J
(52.5)
and is called the error function (we notice that erf oo = 1). Since erf ( — x)
— — erf x, we have finally
T{x y t) =  erf [x/2V( x t)]. (52.6)
Fig. 30 shows a graph of the function erf x. The temperature distribution
becomes more uniform in space in the course of time. This occurs in such a
VO
08
—
l.
<u
04
02
O 02 04 06 08 10 12 14 16 18 2"0
x
Fig. 30
way that any given value of the temperature "moves" proportionally to \/t.
This last result is obviously true. For the problem under consideration is
characterised by only one parameter, the initial temperature difference Tq
between the boundary plane and the remaining space; in the above discussion,
this difference was arbitrarily taken as unity. From the parameters To
and x and variables x and t at our disposal we can form only one dimension
less combination, xj^/{xf)'y hence it is clear that the required temperature
distribution must be given by a function of the form T = To/(a:/\/(xO)'
Let us now consider a case where the surface bounding the medium is a
thermal insulator. That is, there is no heat flux at the plane x = 0, so that
we must have dTjdx = 0. We thus have the following boundary and initial
§52 Thermal conduction in a finite medium 199
conditions :
dT/dx = for x = 0; T = T (x,y,z) for t = 0, * > 0. (52.7)
To find the solution we proceed as in the previous problem. That is, we again
imagine the medium to extend on both sides of the plane x = 0, the initial
temperature distribution being this time symmetrical about the plane. In
other words, we now suppose that To(x, y, z) is an even function of x:
T (  x,y, z) = T (x,y, z). (52.8)
Then dT (x,y, z)\Bx = dT (x,y, z)/dx, and dT Q jdx = for x = 0. It
is evident from symmetry that this condition will continue to be satisfied
for all t.
Repeating the calculations given above, but using (52.8) in place of (52.2),
we have the general solution of the problem in the form
00 00 00
r(w) = i^/J7 r °^>
—oo—oo
{exp[ (x'  xf/4 x t] + exp[ (*' + xfjAxt]} x
exp{  [(/ yf + (z'  zf]l4 x t} dx' dy'dz'. (52.9)
If To is a function of x only, then
T(*>t) = 2^—^ J ^'){exp[(^^/4^] + exp[(^ + ^/4^]}d^'.
(52.10)
Let us now consider problems with boundary conditions of a different type,
which also enable the equation of thermal conduction to be solved in a general
form. Let a heat flux (a given function of time) enter a medium through its
bounding plane x = 0. The boundary and initial conditions are
k8T/8x = q(t) for * = 0; T == for t =  oo, x > 0, (52.11)
where q(t) is a given function.
We first solve an auxiliary problem, in which q(t) = 8(t). It is easy to
see that this problem is physically equivalent to that of the propagation of
heat in an infinite medium from a point source which generates a given
amount of heat. For the boundary condition  kBTjBx = 8(t) for x =
signifies physically that a unit of heat enters through each unit area of the
plane x = at the instant t = 0. In the problem where the condition is
T — 28(x)lpc P for t = 0, an amount of heat
J pc p Tdx = 2
is concentrated on this area at time t = 0; half of this is then propagated in
200 Thermal Conduction in Fluids §52
the positive ^direction, and the other half in the negative ^direction.
Hence it is clear that the solutions of the two problems are identical, and we
find from (51.5) kT(x, t) = V(xH) exp( *a/4tf)
Since the equations are linear, the effects of the heat entering at different
moments are simply additive, and therefore the required general solution of
the equation of thermal conduction with the conditions (52.11) is
t
kT(x, t)= j y_A_ ? ( T ) exp[*2/4 x (*_ T )] dr. (52.12)
—oo
In particular, the temperature on the plane x = varies according to
t
Kr(o < )= jy^ (T)dT  (52  i3)
—oo
Using these results, we can obtain at once the solution of another problem,
in which the temperature T on the plane x = is a given function of time:
T = T Q (t) for x = 0; T = for t =  oo, x > 0. (52.14)
To do so, we notice that, if some function T(x, t) satisfies the equation of
thermal conduction, then so does its derivative dT/dx. Differentiating
(52.12) with respect to x, we obtain
t
8T(x, t) p x i{ r )
— K
= l 2VrXrW a '**Wr)]dr.
8x J 2VI>x(*t) 3 ]
—oo
This function satisfies the equation of thermal conduction and (by (52.11))
its value for x = is q(t); it therefore gives the required solution of the
problem whose conditions are (52.14). Writing T(x, t) instead of — icdT/dx,
and To(t) instead of q(t), we thus have
t
T ^ f > = TTT^ \ TT^Pt* 8 /^'^ ^ (52.15)
—oo
The heat flux q = — kBT/Bx through the bounding plane x = is found by
a simple calculation to be
t
k r dr (T) dr
—00
This formula is the inverse of (52.13).
§52 Thermal conduction in a finite medium 201
The solution is easily obtained for the important problem where the
temperature on the bounding plane x = is a given periodic function of
time: T = Toe~ ib)t for x = 0. It is clear that the temperature distribution
in all space will also depend on the time through a factor e~ M . Since the
onedimensional equation of thermal conduction is formally identical with
the equation (24.3) which determines the motion of a viscous fluid above an
oscillating plane, we can immediately write down the required temperature
distribution by analogy with (24.4) :
T = T exp[*V(W 2 x)] exp{i[xV{(o(2 x )cot]}. (52.17)
We see that the oscillations of the temperature on the bounding surface are
propagated from it as thermal waves which are rapidly damped in the interior
of the medium.
Another kind of thermalconduction problem comprises those concerning
the rate at which the temperature is equalised in a nonuniformly heated
finite body whose surface is maintained in given conditions. To solve these
problems by general methods, we seek a solution of the equation of thermal
conduction in the form T = T n (x, y, z)e~^nt y w ith X n a constant. For the
function T n we have the equation
xAT n = A n 7V (52.18)
This equation, with given boundary conditions, has nonzero solutions only
for certain \ n , its eigenvalues. All the eigenvalues are real and positive,
and the corresponding functions T n (x,y, z) form a complete set of orthogonal
functions. Let the temperature distribution at the initial instant be given by
the function To(x, y, z). Expanding this as a series of functions T n ,
T (x,y,z) = ^c n T n (x,y,z),
we obtain the required solution in the form
T(x, y, z, t) = ^c n T n (x,y, z) exp(X n t). (52.19)
The rate of equalisation of the temperature is evidently determined by the
term corresponding to the smallest A n , which we call Ai. The "equalisation
time" may be defined as t = 1/Ai.
PROBLEMS
Problem 1. Determine the temperature distribution around a spherical surface (of radius
R) whose temperature is a given function T (t) of time.
Solution. The thermalconduction equation for a centrally symmetrical temperature distri
bution is, in spherical coordinates, dTJdt = (x/r)3 8 (rr)/^r 2 . The substitution rT(r, t)
= F(r, t) reduces this to dF/dt = xd^FIdr*, which is the ordinary onedimensional thermal
conduction equation. Hence the required solution can be found at once from (52.15), and is
R(r~R) r T (r)
2r VM J (try
202 Thermal Conduction in Fluids §53
Problem 2. The same as Problem 1, but for the case where the temperature of the spherical
surface is T e~ <wt .
Solution. Similarly to (52.17), we obtain
T = T exp(icot)(Rlr) exp[(li)(rRW(co/2x)].
Problem 3. Determine the temperature equalisation time for a cube of side a whose
surface is (a) maintained at a temperature T = 0, (b) an insulator.
Solution. In case (a) the smallest value of A is given by the following solution of equation
(52.18):
T\ = sm(7Tx/a) s'mfry/a) s\n{iTzJa)
(the origin being at one corner of the cube), when r = 1/A X «= a z J3w*x In case (b) we have
Z 1 ! = cos(7rxJa) (or the same function of y or z), when r = a 2 Jn 2 x
Problem 4. The same as Problem 3, but for a sphere of radius R.
Solution. The smallest value of A is given by the centrally symmetrical solution of (52.18)
Ti = (1/r) sin kr; in case (a), k — nJR, and r = l/ x k* = R 2 Jx** In case (b) k is the
smallest nonzero root of the equation kR = tan kR, whence we find kR = 4493 and
t = 0050 R*/ X 
§53. The similarity law for heat transfer
The processes of heat transfer in a fluid are more complex than those in
solids, because the fluid may be in motion. A heated body immersed in a
moving fluid cools considerably more rapidly than one in a fluid at rest, where
the heat transfer is accomplished only by conduction. The motion of a
nonuniformly heated fluid is called convection.
We shall suppose that the temperature differences in the fluid are so small
that its physical properties may be supposed independent of temperature,
but are at the same time so large that we can neglect in comparison with them
the temperature changes caused by the heat from the energy dissipation
by internal friction (see §55). Then the viscosity term in equation (50.2) may
be omitted, leaving
ar/S/+ vgrad T = X AT, (53.1)
where x = KJpc P is the thermometric conductivity. This equation, together
with the NavierStokes equation and the equation of continuity, completely
determines the convection in the conditions considered.
In what follows we shall be interested only in steady convective flow.f
Then all the time derivatives are zero, and we have the following fundamental
equations :
v.gradr = x AT, (53.2)
(v»grad)v = grad(p/p) + vAv, divv = 0. (53.3)
t In order that the convection should be steady, it is, strictly speaking, necessary that the solid
bodies adjoining the fluid should contain sources of heat which maintain these bodies at a constant
temperature.
§53 The similarity law for heat transfer 203
This system of equations, in which the unknown functions are v, T and pfp,
contains only two constant parameters, v and x Furthermore, the solution
of these equations depends also, through the boundary conditions, on some
characteristic length /, velocity U, and temperature difference T\— To.
The first two of these are given by the dimension of the solid bodies which
appear in the problem and the velocity of the main stream, while the third
is given by the temperature difference between the fluid and these bodies.
In forming dimensionless quantities from the parameters at our disposal,
the question arises of the dimensions to be ascribed to the temperature. To
resolve this, we notice that the temperature is determined by equation (53.2),
which is linear and homogeneous in T. Hence the temperature can be multi
plied by any constant and still satisfy the equations. In other words, the unit
of measurement of temperature can be chosen arbitrarily. The possibility
of this transformation of the temperature can be formally allowed for by
giving it a dimension of its own, unrelated to those of the other quantities.
This can be measured in degrees, the usual unit of temperature.
Thus convection in the abovementioned conditions is characterised by
five parameters, whose dimensions are v = x = cm 2 /sec, U = cm/sec,
/ = cm, 7i— To = deg. From these we can form two independent dimen
sionless combinations. These may be the Reynolds number R = Uljv
and the Prandtl number, defined as
P = vlx. (53.4)
Any other dimensionless combination can be expressed in terms of R and P.f
The Prandtl number is just a constant of the material, and does not depend
on the properties of the flow. For gases it is always of the order of unity.
The value of P for liquids varies more widely. For very viscous liquids, it
may be very large. The following are values of P at 20°C for various
substances :
Air 0733
Water 675
Alcohol 166
Glycerine 7250
Mercury 0044
As in §19, we can now conclude that, in steady convection (of the type
described), the temperature and velocity distributions are of the form
TTq It \ v It \
lwr / (? R 4 u = t (A (53  5)
The dimensionless function which gives the temperature distribution depends
on both R and P as parameters, but the velocity distribution depends only on
R, since it is determined by equations (53.3), which do not involve the con
ductivity. Two convective flows are similar if their Reynolds and Prandtl
numbers are the same.
t The Peclet number is sometimes used; it is defined as UlJx = RP.
204 Thermal Conduction in Fluids §53
The heat transfer between solid bodies and the fluid is usually characterised
by the heat transfer coefficient a, defined by
a = ^/(Tir ), (53.6)
where q is the heat flux density through the surface and T\ To is a charac
teristic temperature difference between the solid body and the fluid. If
the temperature distribution in the fluid is known, the heat transfer coefficient
is easily found by calculating the heat flux density q =  K 8Tjdn at the
boundary of the fluid (the derivative being taken along the normal to the
surface).
The heat transfer coefficient is not a dimensionless quantity. A dimension
less quantity which characterises the heat transfer is what is called the Nusselt
number :f
N = oJ/k. (53.7)
It follows from similarity arguments that, for any given type of convective
flow, the Nusselt number is a definite function of the Reynolds and Prandtl
numbers only:
N=/(R,P). (53.8)
This function is very simple for convection at sufficiently small Reynolds
numbers. These correspond to small velocities. Hence, in the first approxi
mation, we can neglect the velocity term in equation (53.2), so that the
temperature distribution is determined by the equation AT = 0, i.e. the
ordinary equation of steady thermal conduction in a medium at rest. The heat
transfer coefficient can then depend on neither the velocity nor the viscosity
and so we must have simply
N = constant, (53.9)
and in calculating the constant the fluid may be supposed at rest.
PROBLEM
Determine the temperature distribution in a fluid moving in Poiseuille flow along a pipe
of circular crosssection, when the temperature of the walls varies linearly along the pipe.
Solution. The conditions of the flow are the same at every crosssection of the pipe, and
we can look for the temperature distribution in the form T — Az+f(r), where Az is the wall
temperature; we use cylindrical coordinates, with the xaxis along the axis of the pipe.
For the velocity we have, by (17.9), v z = v = 2v m {\ —r 2 jR 2 ), where v m is the mean velocity.
Substituting in (53.2) we find
1 d / d/\ 2VmA
1 d/d/X IVmAV /,yn
rdr\dr) x _ \R/1'
The solution of this equation which is finite for r = and zero for r = R is
VmAr 2
f(r) =
2 X
[(iHGH
f The dimensionless "heat transfer number", denned as Kh ■» ccjpcpU = N/RP, is also used.
§54 Heat transfer in a boundary layer 205
The heat flux density is
q = K [dT/dr] R = \pc v v m RA.
It is independent of the thermal conductivity.
§54. Heat transfer in a boundary layer
The temperature distribution in a fluid at very high Reynolds numbers
exhibits properties similar to those of the velocity distribution. Very large
values of R are equivalent to a very small viscosity. But since the number
P = v fx is not small, the thermometric conductivity x must be supposed
small, as well as v. This corresponds to the fact that, for sufficiently high
velocities, the fluid may be approximately regarded as an ideal fluid, and in
an ideal fluid both internal friction and thermal conduction are absent.
This viewpoint, however, must again be abandoned in a boundary layer,
since neither the boundary condition of no slip nor that of equal temperatures
would be satisfied. In the boundary layer, therefore, there occurs both a
rapid decrease of the velocity and a rapid change of the fluid temperature to a
value equal to the temperature of the solid surface. The boundary layer is
characterised by the presence of large gradients of both velocity and tem
perature.
It is easy to see that, in flow past a heated body (with R large), the
heating of the fluid occurs almost exclusively in the wake, while outside the
wake the fluid temperature does not change. For, when R is large, the pro
cesses of thermal conduction in the main stream are unimportant. Hence the
temperature varies only in the region reached by fluid that has been heated
in the boundary layer. We know (see §34) that the streamlines from
the boundary layer enter the main stream only beyond the line of separation,
where they go into the region of the turbulent wake. From the wake, however,
the streamlines do not emerge at all. Thus the fluid which flows past the
surface of the heated body in the boundary layer goes entirely into the wake
and remains there. We see that the heat becomes distributed through the
regions where the vorticity is nonzero.
In the turbulent region itself, a very considerable exchange of heat occurs,
which is due to the intensive mixing of the fluid characteristic of any turbulent
flow. This mechanism of heat transfer may be called turbulent conduction and
characterised by a coefficient K tuih , in the same way as we introduced the
turbulent viscosity v tmb in §31. The turbulent thermometric conductivity
is defined, in order of magnitude, by the same formula as v turb (31.2):
Xturb ~ l^u.
Thus the processes of heat transfer in laminar and in turbulent flow are
fundamentally different. In the limiting case of very small viscosity and
thermal conductivity, in laminar flow, the processes of heat transfer are
absent, and the fluid temperature is constant at every point in space. In
turbulent flow, however, even in the same limiting case, heat transfer occurs
and rapidly equalises the temperatures in various parts of the stream.
206 Thermal Conduction in Fluids §54
It should be mentioned that, when we speak of the temperature of a fluid in
turbulent motion, we mean the time average of the fluid temperature. The
actual temperature at any point in space undergoes very irregular variations
with time, similar to those of the velocity.
Let us begin by considering heat transfer in a laminar boundary layer.
The equations of motion (39.10) are unaltered. A similar simplification
must now be performed for equation (53.2). Written explicitly, this equation
is (since all quantities are independent of the coordinate z)
dT 8T /d*T d*T\
Vx ~fa +Vy ~dy ~ X \lw + ~df}'
On the righthand side we may neglect the derivative 8 2 TJdx 2 in comparison
with 8 2 T/dy 2 , leaving
BT dT d*T
vx— + v y = x — . (54.1)
dx 8y A dy* v '
By comparing this equation with the first of (39.10) we see that, if the
Prandtl number is of the order of unity, then the order of magnitude 8 of the
thickness of the layer in which the velocity v x decreases and the temperature
T varies will again be given by the formulae obtained in §39, i.e. it will be
inversely proportional to <\/R. The heat flux q =  KdTJdn is equal, in
order of magnitude, to k{T\  T )/ 8. Hence we conclude that q, and therefore
the Nusselt number, are proportional to \/R. The dependence of N on P
is not determined. Thus we have
N = VR/(P). (54.2)
From this it follows, in particular, that the heat transfer coefficient a is
inversely proportional to the square root of the dimension / of the body.
Let us now consider heat transfer in a turbulent boundary layer. Here it
is convenient, as in §42, to take an infinite planeparallel turbulent stream
flowing along an infinite plane surface. The transverse temperature gradient
dT/dy in such a flow can be determined from the same kind of dimensional
argument as we used to find the velocity gradient du/dy. We denote by q
the heat flux density along the jyaxis caused by the temperature gradient.
This flux is a constant (independent of y), like the momentum flux a, and
can likewise be regarded as a given parameter which determines the proper
ties of the flow. Furthermore, we have as parameters also the density p and
the specific heat c p per unit mass. Instead of a we use as parameter v m \
q and c v have the dimensions erg/cm 2 sec = g/sec 3 and erg/g deg = cm 2 /sec 2
deg. The viscosity and thermal conductivity cannot appear explicitly in
dT\dy when R is sufficiently large.
Because of the homogeneity of the equations as regards the temperature,
already mentioned in §53, the temperature can be changed by any factor
without violating the equations. When the temperature is changed in this
way, however, the heat flux must change by the same factor. Hence q and T
§54 Heat transfer in a boundary layer 207
must be proportional. From q, v m p, c v and y we can form only one quantity
proportional to q and having the dimensions deg/cm, namely qjpCp v m y.
Thus we must have dT/dy = ^qjbpc v v m y, where /? is a numerical constant
which must be determined by experiment, j* Hence
T = (Pq/bpc p v m )(\ogy + c). (54.3)
Thus the temperature, like the velocity, varies logarithmically. The constant
of integration c which appears here must be determined from the conditions
in the viscous sublayer, as in the derivation of (42.7). The temperature diff
erence between the fluid at a given point and the wall (which we arbitrarily
take to be at zero temperature) is composed of the temperature change across
the turbulent layer and that across the viscous sublayer. The logarithmic
law (54.3) is determined by only the first of these. Hence, if we write (54.3)
in the form T = (Pqlbpc p v m )[log(yvJv) + constant], including in the argument
of the logarithm a factor equal to the thickness yo, then the constant (multi
plied by the coefficient in parentheses) must be the change in temperature
across the viscous sublayer. This change, of course, depends on the coeffi
cients v and x also. Since the constant is dimensionless, it must be some
function of P, which is the only dimensionless combination of the quantities
v, x, p, v# and c v (q cannot appear, since T must be proportional to q, which
already occurs in the coefficient). Thus we find the temperature distribution
to be
T = (pqlbpc p v*)[log(vvJv) +/(P)]. (54.4)
Using this formula, we can calculate the heat transfer for turbulent flow in a
pipe, along a flat plate, etc. We shall not pause to do this here.
PROBLEMS
Problem 1 . Determine the limiting form of the dependence of the Nusselt number on the
Prandtl number in a laminar boundary layer when P and R are large.
Solution. For large P, the distance 8' over which the temperature changes is small
compared with the thickness 8 of the layer in which the velocity v x diminishes. 8' may be
called the thickness of the temperature boundary layer. The order of magnitude of 8' may
be obtained from an estimate of the terms in equation (54.1). Over the distance from y —
to y fj 8', the temperature varies by an amount of the order of the total temperature diff
erence T r — T between the fluid and the solid body, while the velocity v x varies over this
distance by an amount of the order of C/S'/S (since the total change, of the order of U, occurs
over a distance 8). Hence, for y <>' 8', the terms in equation (54.1) are, in order of magnitude,
x d2T/dyZ ~ x (Ti 7o)/S' 2 and v x dT/dx ~ C/S'(Ti T )//S.
If the two expressions are comparable, we have S' 3 "~ x/S/t7. Substituting 8 ~ l\ VR,
we obtain 8' ~ Z/Rip* ^/ S/Pi. Thus, for large P, the thickness of the temperature boundary
layer decreases, relative to that of the velocity boundary layer, inversely as the cube root of P.
f Here b is the constant appearing in the logarithmic velocity profile (42.4). With this definition,
P is the ratio vtart>/xturb> where rturb and xturb are the coefficients in q = pc P xturbdT/dy,
a = pvtnrbdtt/dy. From simultaneous determinations of the velocity and temperature profiles in
pipes and in flow along flat plates,/? is found to be about 07. We should mention that similar measure
ments in the turbulent wake behind a heated body give a value of about 0 5 for the ratio vturb/xturb in
a free turbulent flow.
208 Thermal Conduction in Fluids §54
The heat flux q = — KdTJdy ~ k(T x  T )jo", and the required limiting law of heat transfer
is found to bef
N = constant x R*P*.
Problem 2. Determine the limiting form of the function /(P), in the logarithmic tempera
ture distribution (54.4), for large values of P.
Solution. According to what was said in §42, the transverse velocity in the viscous sub
layer is of the order of v*(yfy ) 2 , while the scale of the turbulence is of the order of y*fy .
The turbulent thermometric conductivity x turb is therefore of the order of
»*.yoCv/yo) 4 ~ v(ylyo) 4
(where we have used the relation (42.5)); xturb is comparable in magnitude with the ordinary
coefficient x at distances of the order of y r ~ y oP  *. Since x turb increases very rapidly with y,
it is clear that most of the temperature change in the viscous sublayer occurs over distances
from the wall of the order of y u and may be supposed proportional to y lf being in order of
magnitude qyj k ~ qy f kP* ~ qP*/pc P v*. Comparing with formula (54.4), we see that the
function /(P) is a numerical constant times P*.J
Problem 3. Determine the temperature differences T^ in a nonuniformly heated turbulent
fluid over distances A which are small compared with the external scale of the turbulence
(A. M. Obukhov 1949).
Solution. The equalisation of temperature in a nonuniformly heated turbulent fluid
occurs similarly to the dissipation of mechanical energy. Turbulent eddies of size A^> A
(where A is the internal scale of the turbulence) lead to an equalisation of temperature by
purely mechanical mixing of fluid particles which are at different temperatures. Consider
able true temperature gradients in regions of size A ~ A , on the other hand, are equalised
by dissipative thermal conduction.
The dissipation by thermal conduction (increase of entropy) is determined by the quantity
x(grad TflT* (see (49.6)); supposing the turbulent fluctuations of temperature to be rela
tively small, we can replace T 2 in the denominator by a constant, the square of the mean
temperature. According to the method described in §32 (see the first footnote to that section),
we write Xturb (7a/A) 2 = constant. Substituting v t urb,A ~ »Wb,A ~ ~^x> *>A ~ (**)* (see
(32.1)), we find the required relation to be T^ ~ A*. Thus for A^> A the temperature fluc
tuations, like the velocity fluctuations, are proportional to the cube root of the distance.
At distances A <^ A , however, by the same arguments as for the velocity, the differences T^
are simply proportional to A.
Problem 4. Derive a relation between the local correlation functions
B TT = (T 2  7i)2, B iTT = (v 2i v u )(T 2  Ti)2
in a nonuniformly heated turbulent flow (A. M. Yaglom 1949).
Solution. The calculations are similar to those used in deriving formula (33.18). From
t For the values of the thermal conductivity actually found, the Prandtl number does not reach
the values for which this limiting law holds. Such laws can, however, be applied to convective diffu
sion; this obeys the same equations as convective heat transfer, but with the temperature replaced by
the concentration of the solute, and the heat flux by the flux of solute, the "diffusion Prandtl number"
being defined as Po = v/D, where D is the diffusion coefficient. For example, for solutions in water
and similar liquids, Pj> reaches values of the order of 10 3 , while for very viscous solvents it is 10* or
more.
J The calculation of the constant in this formula for various particular cases is facilitated by the fact
that, by virtue of the inequality S' <^ S, we need take only the first terms of an expansion, in powers of
y, of the fluid velocity components in integrating equation (54.1) across the temperature boundary
layer. Calculations for convective diffusion in various particular cases are given by V. G. Levich,
Physicochemical Hydrodynamics {Fizikokhimicheskaya gidrodinamika) , Moscow 1952.
§55 Heating of a body in a moving fluid 209
the equations
dT dT dvi n
— + vi — = x&T, — =
dt dxi oxi
we find
d
(7i T 2 ) = 2— (wT! T 2 ) + 2 x Ai(T 1 T 2 ).
Ot OX\i
On the lefthand side we put r = r 2 — r u and on the right we express the mean values in terms
of the correlation functions, using the homogeneity and isotropy of the flow:
a— l d
— T 2 =  — BittxAiBtt
ot 2 ox\i
Writing B t TT = «< B t tt and changing to derivatives with respect to r, we obtain an equation
which, on integration over r, gives the required relation
B r TT — 2xdBTTldr = —&<(>>
where
4> = d(T*)\dt = 8(TT) 2 ldt.
Using the results of Problem 3, we then find that, for r^> A , B t tt = — ir<f>, while for
r <^ A we have Btt = r 2 <f>/9x
§55. Heating of a body in a moving fluid
A thermometer immersed in a fluid at rest indicates a temperature equal to
that of the fluid. If the fluid is in motion, however, the thermometer indicates
a somewhat higher temperature. This is because the fluid brought to rest
at the surface of the thermometer is heated by internal friction.
The general problem may be formulated as follows. A body of arbitrary
shape is immersed in a moving fluid ; thermal equilibrium is established after
a sufficient length of time, and it is required to determine the temperature
difference T\ — To then existing between the body and the fluid.
The solution of this problem is given by equation (50.2), in which, however,
we cannot now neglect the term containing the viscosity as we did in (53.1);
it is this term which is responsible for the effect under consideration. Thus
we have for a steady state
vgrad r = xA r + _(_! + _± . (55.1)
ZCp \ OXjc OX{ ]
This must be supplemented by the equations of motion (53.3) of the fluid
itself and also, strictly speaking, by the equation of thermal conduction in the
body. In the limiting case where the body has a sufficiently small thermal
conductivity, we can neglect the latter and suppose the temperature at any
point on the surface of the body to be simply equal to the fluid temperature
at that point, obtained by solving equation (55.1) with the boundary condition
210 Thermal Conduction in Fluids §55
dTjdn = 0, i.e. the condition that there is no heat flux through the surface of
the body. In the opposite limiting case where the body has a sufficiently
large thermal conductivity, we can use the approximate condition that the
temperature should be the same at every point of its surface; the derivative
dTjdn will not then in general vanish over the whole surface, and we must
require only that the total heat flux through the surface of the body (i.e. the
integral of BT/dn over the surface) should be zero. In both these limiting cases
the thermal conductivity of the body does not appear explicitly in the solution
of the problem, and we shall suppose in what follows that one of these cases
holds.f
Equations (55.1) and (53.3) contain the constant parameters x> v and c v ,
and their solutions involve also the dimension / of the body and the velocity
U of the main stream. (The temperature difference TiT is not now an
arbitrary parameter, but must itself be determined by solving the equations.)
From these parameters we can construct two independent dimensionless
quantities, which we take to be R and P. Then we can say that the required
temperature difference 7i  To is equal to some quantity having the dimensions
of temperature (which we take to be U 2 jc v ), multiplied by a function of R and
P:
r 1 r = (^/c 3 ,)/(R,p). (55.2)
It is easy to determine the form of this function for very small Reynolds
numbers, i.e. for sufficiently small velocities U. In this case the term
V'grad T in (55.1) is small compared with xAT, so that this equation be
comes
XAT= 
ZCp
The temperature and velocity vary considerably over distances of the order
of /. Hence an estimate of the two sides of equation (55.3) gives x(T\ T )/l 2
~vU 2 /c p l 2 , or TiT ~ vU 2 Jxc p . Thus we conclude that, for small R,
7i  To = constant x P U 2 /c p , (55.4)
where the numerical constant depends on the shape of the body. It should
be noticed that the temperature difference is proportional to the square of
the velocity U.
Some general conclusions concerning the form of the function /(P, R) in
(55.2) can be drawn in the opposite limiting case of large R, when the velocity
and the temperature vary only in a narrow boundary layer. Let 8 and S' be
the distances over which the velocity and temperature respectively vary; 8
and 8' differ by a factor depending on P. The amount of heat evolved in
unit area of the boundary layer in unit time owing to the viscosity of the fluid
t I. A. Kibel' has obtained an exact solution for the rotation of a heated disk in a viscous fluid, simi
lar to the solution given in §23 for a constant temperature; see Prikladnaya matematika i mekhanika 11,
611, 1947.
§55 Heating of a body in a moving fluid 211
is the integral of \vp{dvijdx]c\dvjcjdxi) 2 over the thickness of the layer
(see (16.3)). This integral is of the order of vp(U 2 jB 2 )8 — vpU 2 j8. The same
amount of heat must be lost to the body, and it is therefore equal to the heat
flux q =» — KdTjdn ~ x c vp{Ti — 7o)/o". Comparing the two expressions, we
find
T 1 T = (U 2 lc p )f(P). (55.5)
Thus, in this case, the function /is independent of R, but its dependence on
P remains undetermined.
PROBLEMS
Problem 1. Determine the temperature distribution in a fluid moving in Poiseuille flow
in a pipe of circular crosssection whose walls are maintained at a constant temperature T .
Solution. In cylindrical coordinates, with the araxis along the axis of the pipe, we have
v z — v = 2vm[l — (r/R)*\, where v m is the mean velocity of the flow. Substitution in (55.3)
gives the equation
1 d / dT\ \6v m 2 „
1 r I = r 2 .
r dr \ dr J R* x c v
The solution finite at r = and equal to T for r = R is
rr„ = ^[i(L)] 4 .
Problem 2. Determine the temperature difference between a solid sphere and a fluid
moving past it at small Reynolds numbers. The thermal conductivity of the sphere is
supposed large.
Solution. We take spherical coordinates r, 8, <f>, with the origin at the centre of the sphere
and the polar axis in the direction of the velocity of the main stream. Calculating the com
ponents of the tensor dvildxk+dvk/dxi by means of formulae (15.17) and (20.9), we obtain
equation (55.3) in the form
1 d / „dT\ 1 d /
1 (
r* 8r\
r z I + ( sin0
dr J r*smd dd\ 86/
=  A(Rlry [cos20{3  6(i?/r)2 + 2(i?/r)4} + (R/r)*] ,
where A = 9« a P/4c„. We look for T(r, 0) in the form T =f(r) cos 2 +g(r), and, separating
the part which depends on 9, find two equations for /and g:
r 2 f" + 2rf6f = A[3{R/r) 2 6(R/ry+2(Rlr)%
r 2 g" + Irg' + 2/ =  A(R/r)*.
From the first we obtain
/ = A[^R/r) 2 + (R/rYMRIr) 6 ] + ci(Rlrf;
the term of the form constant X r 2 is omitted, since it does not vanish at infinity. The second
equation then gives
g = lAB(Rlr) 2 + i(RlrY+MRIr) 6 ]MR/rf+c 2 Rlr+C3.
The constants c u c 2 , c a are determined from the conditions
T= constant and f (dT/dr)r* sin0d0 =
212 Thermal Conduction in Fluids §56
for r = R, which are equivalent to/(JR) = and g'(R)+\f'(R) = 0; also T = T at infinity.
Thus c x = —5.4/3, c 2 = 2.4/3, c 8 = T . The temperature difference between T t = T(i?)
and T is found to be T a — T = 5u 2 P/8c P . It may be noted that the temperature distribution
obtained actually satisfies the condition dT/dr — for r = R, i.e. / '(R) = £'(.R) = 0.
Hence it is also the solution of the same problem for a sphere of small thermal conductivity.
§56. Free convection
We have seen in §3 that, if there is mechanical equilibrium in a fluid in a
gravitational field, the temperature distribution can depend only on the alti
tude z: T = T(z). If the temperature distribution does not satisfy this
condition, but is a function of the other coordinates also, then mechanical
equilibrium in the fluid is not possible. Furthermore, even if T = T(z),
mechanical equilibrium may still be impossible if the vertical temperature
gradient is directed downwards and its magnitude exceeds a certain value (§4).
The absence of mechanical equilibrium results in the appearance of internal
currents in the fluid, which tend to mix the fluid and bring it to a constant
temperature. Such motion in a gravitational field is called free convection.
Let us derive the equations describing this convection. We shall suppose
the fluid incompressible. This means that the pressure is supposed to vary
only slightly through the fluid, so that the density change due to changes in
pressure may be neglected. For example, in the atmosphere, where the pres
sure varies with height, this assumption means that we shall not consider
columns of air of great height, in which the density varies considerably over
the height of the column. The density change due to the nonuniform heating
of the fluid, of course, can not be neglected ; it results in the forces which
bring about the convection.
We write the variable temperature T(x, y, z, t) in the form T = To + T',
where To is some constant mean temperature from which the variation T' is
reckoned. We shall suppose that 7" is small compared with To.
We write the fluid density also in the form p = p + p, with po a constant.
Since the temperature variation 7" is small, the resulting density change p
is also small, and we can write
p' = (8pol8T) p T f = pojSr. (56.1)
Here /? = —(l/p)dpJ8T is the thermalexpansion coefficient of the fluid.
In the pressure p = po+p', Po is not constant. It is the pressure cor
responding to mechanical equilibrium, when the temperature and density are
constant and equal to To and po respectively. It varies with height according
to the hydrostatic equation
Po = Pog*^ + constant. (56.2)
We start by transforming the NavierStokes equation, which has, in the
presence of a gravitational field, the form
8v/d* + (v»grad)v = (l//>)gradp + vAv+g;
this is obtained by adding the force g per unit mass to the righthand side
§56 Free convection 213
of equation (15.7). We now substitute p = po+p\ p = po+p; to the first
order of small quantities, we have
grad^> grad/> gradp' grad^>
—  — = + p ' t
P po po po*
or, substituting (56.1) and (56.2),
grad/> grad/>'
  = g +   + gT%
P po
With this expression, the NavierStokes equation gives
Sv/^+(v.grad)v= (l//>)grad/>' + vAvj8rg, (56.3)
where the suffix has been dropped from p . In the thermal conduction equa
tion (50.2), the viscosity term can be shown to be small in free convection
compared with the other terms, and may therefore be omitted. We thus
obtain
dT'/dt+vgrad T = X AT*. (56.4)
Equations (56.3) and (56.4), together with the equation of continuity
div v = 0, form a complete system of equations governing free convection.
For steady flow, the equations of convection become
(v.grad)v= (l//>)grad£'j8rg + vAv, (56.5)
v.gradT' = X AT', (56.6)
divv = 0. (56.7)
This system of five equations for the unknown functions v, p'[p and V
contains three parameters, v, x and fig. Moreover, the solution will involve
a characteristic length / and the temperature difference Ti  T between the
solid body and the fluid at a great distance. There is here no characteristic
velocity, since there is no flow due to external forces, and the whole motion
of the fluid is due to its nonuniform heating.
Thus steady free convection in a gravitational field is characterised by
five parameters, which have the following dimensions: x = v = cm 2 /sec,
Ti —Tq = deg, / = cm, fig = cm/sec 2 deg. From these we can form two
independent dimensionless quantities, which we take to be the Prandtl
number P = v/x and the Grashof number
G = ^(T 1 T )I^. (56.8)
The similarity law for free convection is therefore
v = (v//)f(r//, G), T = (Ti  7o)/(r//, P, G). (56.9)
Two flows are similar if their Prandtl and Grashof numbers are the same.
Convective heat transfer caused by gravity is again characterised by the
214 Thermal Conduction in Fluids §56
Nusselt number, which is now a function of P and G only:
N=/(P,G). (56.10)
The value of the Grashof number is an important characteristic of con
vective flow. When G is sufficiently small, the free convection is unimportant
in the heat transfer in the fluid, which is then due mainly to ordinary con
duction.
Convective flow may be either laminar or turbulent. There is no Reynolds
number for free convection (since there is no characteristic velocity para
meter), and the onset of turbulence is determined by the Grashof number :
the convection becomes turbulent when G is very large.
A very curious case of convection is the flow which occurs in a fluid between
two infinite horizontal planes at different temperatures, that of the lower plane
(T2) being greater than that of the upper plane (Ti). If the temperature
difference T% — T\ is small, the fluid remains at rest and there is pure thermal
conduction, the fluid temperature and density being functions only of the
vertical coordinate z; the density increases upward. If the difference T2—T1
exceeds a certain critical value, however, which depends on the distance /
between the planes, such a state becomes unstable and steady convection
occurs. The onset of instability can be determined theoretically (see Problem
5). The critical value of the difference T%— T\ appears as a factor in the
product
GP = ^Z3(r 2 Ti)/v X . (56.11)
In a layer of fluid between two solid planes at constant temperatures, con
vection must occur if GP > 1710. If the upper surface is free, but still at a
constant temperature, then convection occurs for GP > HOO.f
The convective flow which occurs is somewhat unusual. Since the fluid is
unbounded in the horizontal plane, it is evident that the flow must be periodic
in that plane. In other words, the space between the bounding planes must be
divided into similar right prisms in each of which the fluid moves in a similar
way. The horizontal crosssections of these prisms form a network in the
horizontal plane. The theoretical determination of the nature of this network
is very difficult, but experimental results seem to indicate that there is a
hexagonal pattern with cells in the form of hexagonal prisms, the fluid moving
up in the middle and down at the edges, or else vice versa.
For very large values of G, the steady convection in turn becomes unstable ;
turbulence sets in for G ~ 50,000.
Another similar case of instability is that of convection in a vertical
cylindrical pipe along which a constant temperature gradient is maintained.
t These conditions (for a given difference T 2 — T t ) are always fulfilled if / is sufficiently large. To
avoid misunderstanding, we should mention that we are speaking here of values of / for which the
variation in the fluid density under the action of gravity is unimportant. Hence the above criteria
cannot.be applied to gas columns of great height. In this case we have to use the criterion derived in
§4, from which we see that convection need not occur for a column of any height if the temperature
gradient is small enough.
§56 Free convection 215
Here again there is a critical value of the product GP beyond which the fluid
at rest is unstable; see Problem 6.
PROBLEMS
Problem 1. Determine the Nusselt number for free convection on a flat vertical plate.
It is assumed that the velocity and the temperature difference T = T—T (where T is
the fluid temperature at infinity) are appreciably different from zero only in a thin boundary
layer adjoining the surface of the plate (K. PohlhaUsen).
Solution. We take the origin on the lower edge of the plate, the *axis vertical, and the
yaxis perpendicular to the plate. The pressure in the boundary layer does not vary along the
}>axis (cf. §39), and therefore is everywhere equal to the hydrostatic pressure p (x), i.e.
p' = 0. With the usual accuracy of boundarylayer theory, equations (56.5)(56.7) become
a y =^ + ^r„), (i)
BT BT B*T
. \ Vy = ^ —
Bx By By
+ *«r£: = xzi. ( 2 )
Bv x Bv y
with the boundary conditions v x = v y = and T = T± for y = (T x being the temperature
of the plate), v x = and T = T for y = oo. These equations can be converted into ordinary
differential equations by introducing as the independent variable
€ = Cyjx\ C = \MTi ?o)/4v2]*. (4)
We put
v x = AvCW^'^l TT = (T 1  r o )0(£). (5)
Then (3) gives v y = vCxh({<f>'3<f>), and (1) and (2) give equations for <f> and 0:
41" + 3cfxf>"24'2 + d = 0, 6" + 3P<£0' = 0, (6)
with the boundary conditions #0) = f (0) = 0, 6(0) = 1, f(oo) = 0, «(oo) = 0. It follows
from (4) and (5) that the thickness of the boundary layer is of the order S ~ x i JC. The con
dition for the solution to be valid is therefore S < / (where I is the height of the plate), or
G* ^> 1. The total heat flux per unit area of the plate is
The Nusselt number is N =/(P)G*, where the function /(P) is determined by solving the
equations (6).
Problem 2. A hot turbulent submerged jet of gas is bent round by a gravitational field : find
its shape (G. N. Abramovich 1938).
Solution. Let T be some mean value (over the crosssection of the jet) of the temperature
difference between the jet and the surrounding gas, u some mean velocity of the gas in the
jet, and / the distance along the jet from its point of entry; I is supposed large compared with
the dimensions of the aperture by which the jet enters. The condition of constant heat flux
Q along the jet is Q ~ pc^TuR^ = constant and, since the radius of a turbulent jet is pro
portional to / (cf. §35), we have
T'ul 2 = constant ~ Q/pCp] (1)
216 Thermal Conduction in Fluids §56
we notice that, in the absence of the gravitational field, u f* \jl (see (35.3)) and it then follows
from (1) that T ~ 1/Z.
The momentum flux vector through the crosssection of the jet is proportional to pu 2 R 2 n
~ gu 2 l 2 n, where n is a unit vector along the jet. Its horizontal component is constant along
the jet:
m 2 / 2 cos 6 = constant, (2)
id the horizontal, while the chang
et. This force is proportional to
ppgT'R* ~ ppgT'l* ~ feQlcp.
where 8 is the angle between n and the horizontal, while the change in the vertical component
is due to the "lift force" on the jet. This force is proportional to
Hence we have
d(Z 2 w 2 sin 0)/dZ  PgQIpcpU. (3)
It then follows from (2) that d(tan 0)/dl = constant X / cos* 0, whence we obtain finally
/
dd
= constant x I 2 , (4)
cos 5/2
where O gives the direction of the emergent jet.
In particular, if does not vary appreciably along the jet, (4) gives 0— O = constant X/ 2 .
This means that the jet is a cubical parabola, in which the deviation d from a straight line is
d = constant X I 3 .
Problem 3. A turbulent jet of heated gas (i.e. one with a large Grashof number) rises from
a fixed hot body. Determine the variation of the velocity and temperature in the jet with
height (Ya. B. Zel'dovich 1937).
Solution. As in the preceding case, the radius of the jet is proportional to the distance
from its source, and we have, analogously to (1) of Problem 2, T'uz 2 = constant, and instead
of (3) d(z 2 u 2 )Jdz = constant/w, where z is the height above the body, supposed large compared
with the dimension of the body. Integrating, we find u ~ z~*, and for the temperature
r ~z s '\
Problem 4. The same as Problem 3, but for a laminar convective jet rising freely (Ya.
B. Zel'dovich 1937).
Solution. Together with the relation T'uR 2 = constant, which expresses the constancy
of the heat flux, we have u 2 fz ~ vu/R 2 ~ PgT', which follows from equation (56.5). From
these relations we find the following variation of the radius, velocity and temperature with
height: R ~ ^z, u = constant, T ~ 1/z. It may be noticed that the number G ~ T'R 3
~ \/z, i.e. increases with height, and the jet must therefore become turbulent at a certain
altitude.
Problem 5. Derive the equations governing the onset of steady convection between two
horizontal planes maintained at given temperatures (Rayleigh 1916).
Solution. A perturbation proportional to e~ iu>t is applied to a fluid at rest with a constant
vertical temperature gradient dT/dz = —A < 0. The state of rest is unstable if there is any
possible value of to whose imaginary part is positive. Hence the onset of instability is deter
mined by the appearance of a solution for which the imaginary part of w is zero. In this case
we are concerned with the appearance of steady convection as a result of instability; hence
we must seek solutions for which the real part of w is also zero, that is, solutions independent
of time.
In equations (56.5)(56.7), the velocity v of the perturbing motion and the resulting pressure
variation p' are small quantities. We write the temperature as T = —Az+r, where the
§56 Free convection 217
perturbation t is small ; we suppose the pressure variation resulting from the constant tem
perature gradient to be included in p . Then we find, omitting secondorder terms,
vAv = gjcad(p'lp)+pTg,
xAr = Av Zt divv = 0. * '
Eliminating v and p'/p, we obtain an equation for t:
where y = V$gA\vx — GP, and Z is the distance between the planes.
The boundary conditions on equations (1) at a solid surface are t = 0, v z — 0, 8v z \8z — 0.
The last of these follows from the equation of continuity, since we must have v x = v y =
for all x and y. By the second equation (1), the conditions on v z can be replaced by conditions
on higher derivatives of t, c 2 being replaced by 8 2 rj8z 2 .
We look for t in the form e ik ' T f(z), where k is a vector in the xyplane, and obtain for f(z)
the equation
d 2 \3 yk 2
1 d* \ 3 y&
The general solution of this equation is a linear combination of the functions cosh((iz/l)
and sinh^z/l), where n 2 = k 2 l 2 —y*(klfty\ with the three different values of ^/\. The
coefficients are determined by the boundary conditions, which lead to a system of algebraic
equations; the compatibility condition then determines the function kl(y). The inverse
function y = y(kl) has a minimum for some value of kl; the corresponding y = GP deter
mines the required criterion for the appearance of instability, and the value of k determines
the periodicity in the xyplane, but not the symmetry, of the resulting motion, f
Problem 6. Determine the onset of steady convection in a fluid at rest in a vertical cylin
drical pipe along which a constant temperature gradient is maintained (G. A. Ostroumov
1946).
Solution. We seek a solution of the equations (1) of Problem 5 in which the convective
velocity v is everywhere parallel to the axis of the pipe (the .saxis), and the flow pattern does
not vary along this axis, i.e. v z = v, r and dp'ldz depend only on the coordinates xand y.
Then the equations become Bp'Jdx = 0, dp' I By = 0, vA 2 v = —PgT+(l[ P )dp'ldz, xA 8 t
= — Av, where A 2 = d 2 /8x 2 + 8 2 j8y 2 . The first two equations show that dp'fdz = constant,
and, eliminating t from the other equations, we have
where we have again put y = AR^g/xv = GP, and R is the radius of the pipe. At the surface
of the pipe we must have v = and the heat flux continuous. Moreover, the total mass flux
through a crosssection of the pipe must be zero.
Equation (1) has solutions of the form Jn(kr) cos rt(f> and I n (kr) cos n<f>, where J n and I n
are Bessel functions of real and imaginary argument respectively, r and <j> are polar coordi
nates in the crosssection, and kR = y*. The onset of convection corresponds to the solution
for which y is least. It is found that this is the solution with n = 1 :
v = vocostUiikry^ktyhikrViikR)],
r = voiv&l^costtMkrWkiq + hikrViikR)].
f A detailed account of the calculations is given by A. Pellew and R. V. Southwell, Proceedings
of the Royal Society A176, 312, 1940.
218 Thermal Conduction in Fluids §56
The pressure gradient dp'Jdz does not appear. The condition v = for r — R is satisfied
identically, and the total mass flux through the crosssection of the pipe is zero. In the
limiting case of thermally insulating walls, we must have also drjdr = for r = R, or
JojkR) IojkR) _ 2
j!(kR) + h(kR) kR
The smallest root of this equation gives the required critical value of y = (kR)* = 674.
In the opposite limiting case of walls of infinite thermal conductivity, we must have t =
for r = R; thenJi(kR) = 0, whence the critical value is y = 215*8.f
t For a more detailed discussion see G. A. Ostroumov, Free Convection in a Confined Medium
(Svobodnaya konvektsiya v usloviyakh vnutrennei zadachi), Moscow 1952.
CHAPTER VI
DIFFUSION
§57. The equations of fluid dynamics for a mixture of fluids
Throughout the above discussion it has been assumed that the fluid is
completely homogeneous. If we are concerned with a mixture of fluids
whose composition is different at different points, then the equations of
fluid dynamics are considerably modified.
We shall discuss here only mixtures with two components. The com
position of the mixture is described by the concentration c, defined as the
ratio of the mass of one component to the total mass of the fluid in a given
volume element.
In the course of time, the distribution of the concentration through the
fluid will in general change. This change occurs in two ways. Firstly, when
there is macroscopic motion of the fluid, any given small portion of it moves
as a whole, its composition remaining unchanged. This results in a purely
mechanical mixing of the fluid; although the composition of each moving
portion of it is unchanged, the concentration of the fluid at any point in space
varies with time. If we ignore any processes of thermal conduction and inter
nal friction which may also be taking place, this change in concentration is a
thermodynamically reversible process, and does not result in the dissipation
of energy.
Secondly, a change in composition can occur by the molecular transfer of
the components from one part of the fluid to another. The equalisation of the
concentration by this direct change of composition of every small portion of
fluid is called diffusion. Diffusion is an irreversible process, and is, like
thermal conduction and viscosity, one of the sources of energy dissipation in a
mixture of fluids.
We denote by p the total density of the fluid. The equation of continuity
for the total mass of the fluid is, as before,
8 P l8t + div(pv) = 0. (57.1)
It signifies that the total mass of fluid in any volume can vary only by the
movement of fluid into or out of that volume. It must be emphasised that,
strictly speaking, the concept of velocity itself must be redefined for a mixture
of fluids. By writing the equation of continuity in the form (57.1), we have
defined the velocity, as before, as the total momentum of unit mass of fluid.
The NavierStokes equation (15.5) is also unchanged. We shall now derive
the remaining equations of fluid dynamics for a mixture of fluids.
In the absence of diffusion, the composition of any given fluid element
would remain unchanged as it moved about. This means that the total
219
220 Diffusion §57
derivative dcjdt would be zero, i.e. the equation dc/dt = dc/dt+vgrad c =
would hold. This equation can be written, using (57.1), as
d(pc)ldt + div(pcv) = 0,
i.e. as an equation of continuity for one of the components of the mixture
(pc being the mass of that component in unit volume). In the integral form
— pcdV = — (b pcvdf
it shows that the rate of change of the amount of this component in any
volume is equal to the amount of the component transported through the
surface of that volume by the motion of the fluid.
When diffusion occurs, besides the flux pcv of the component in question
as it moves with the fluid, there is another flux which results in the transfer
of the components even when the fluid as a whole is at rest. Let i be the
density of this diffusion flux, i.e. the amount of the component transported
by diffusion through unit area in unit time.f Then we have for the rate of
change of the amount of the component in any volume
— pcdV = — (ppcv'df— ()i»df,
or, in differential form,
d{pc)\dt =  div(pcv)  div i. (57.2)
Using (57.1), we can rewrite this "equation of continuity" for one component
in the form
p(dcjdt+v grade) = divi. (57.3)
To derive another equation, we repeat the arguments given in §49, bearing
in mind that the thermodynamic quantities for the fluid are now functions of
the concentration also. In calculating the derivative 8(%pv 2 + pe)jdt (in §49)
by means of the equations of motion, we had to transform the terms pdejdt
and — vgradp. This transformation must now be modified, because the
thermodynamic identities for the energy and the heat function now contain an
additional term involving the differential of the concentration:
dc = Tds+(plpZ)dp + fj,dc,
dw = Tds + (l[p)dp+fidc,
f The sum of the flux densities for the two components must be pv. If the flux density for one
component is pev+i, that for the other component is therefore p(l — c)v— i.
§57 The equations of fluid dynamics for a mixture of fluids 221
where p. is an appropriately denned chemical potential of the mixture.f
Accordingly, an additional term p/idc/dt appears in the derivative pde/dt.
Writing the second thermodynamic relation in the form
dp = pdw—pTds—pfxdc,
we see that the term — vgradp will contain an additional term ppv grad c.
Thus we must add pp{dcJdt + \ grade) to the expression (49.3). By
equation (57.3), this can be written p. div i. The result is
d
—(ipv 2 +pe) = — div\pv(lv 2 + w)— v»o' + q] +
8t
(8s \ dvt
+pT\ — + vgrads)  a'ik h divqju, divi. (57.4)
\ dt 1 dxic
We have replaced  k grad T by a heat flux q, which may depend not only
on the temperature gradient but also on the concentration gradient (see the
next section). The sum of the last two terms on the right can be written
divq— jLtdivi = div(q— jui)+igrad^.
The expression p\(%v 2 + w) — v«a' + q which is the operand of the diver
gence operator in (57.4) is, by the definition of q, the total energy flux in
the fluid. The first term is the reversible energy flux, due simply to the
movement of the fluid as a whole, while the sum — vo' + q is the irreversible
flux. When there is no macroscopic motion, the viscosity flux v«o' is zero,
and the thermal flux is simply q.
The equation of conservation of energy is
8
—Upv 2 +pe)= div[pv(t; 2 + ro)v.o' + q]. (57.5)
Subtracting from (57.4), we obtain the required equation
8s _ , \ , 3©<
8xjc
which is a generalisation of (49.4).
(8s \ dvt
P T \ ~Z + v 'S rad * = a ' ik ~o div(qju)i.grad/A, (57.6)
\ ot J dxic
t It is known from thermodynamics that, for a mixture of two substances, the thermodynamic
identity is
de = Tds— pdV+fiidni+p,2dn2,
where n lt n^ are the numbers of particles of the two substances in 1 g of the mixture, and (j. lt ^ are
the chemical potentials of the substances. The numbers n v n^ satisfy the relation « 1 OT 1 +n 2 OT 2 = 1,
where m 1 and n^ are the masses of the two kinds of particle. If we introduce as a variable the
concentration c = n^m x , we have
de= TdspdV+(^^)dc.
\m\ mil
Comparing this with the relation given in the text, we see that the chemical potential fi is related
to fa and /ig by
_ pi H>2
mi 7»2
222 Diffusion §58
We have thus obtained a complete system of equations of fluid mechanics
for a mixture of fluids. The number of equations in this system is one more
than for a single fluid, since there is one more unknown function, namely the
concentration. The equations are the equation of continuity (57.1), the
NavierStokes equations, the "equation of continuity" (57.2) for one com
ponent, and equation (57.6), which determines the change in entropy.
It must be noticed that equations (57.2) and (57.6) as they stand determine
only the form of the corresponding equations of fluid dynamics, since they
involve the undetermined fluxes i and q. These equations become determi
nate only when i and q are replaced by expressions in terms of the gradients
of concentration and temperature. The corresponding expressions will be
obtained in §58.
For the rate of change of the total entropy of the fluid, a calculation entirely
similar to that of §49, but using (57.6) in place of (49.4), gives the result
1 J„dF   J ^yV J^dr + ..., (57.7)
where we have omitted, for brevity, the viscosity terms.
§58. Coefficients of mass transfer and thermal diffusion
The diffusion flux i and the heat flux q are due to the presence of con
centration and temperature gradients in the fluid. It should not be thought,
however, that i depends only on the concentration gradient and q only on the
temperature gradient. On the contrary, each of these fluxes depends, in
general, on both gradients.
If the concentration and temperature gradients are small, we can suppose
that i and q are linear functions of grad fi and grad T.f Accordingly,
we write i and q as
i= agrad/AjSgradT, q= S grad [My grad T+fxi.
There is a simple relation between the coefficients /? and 8, which is a
consequence of a symmetry principle for the kinetic coefficients. This symmetry
principle is as follows.!
Let us consider some closed system, and let xi, X2, ... be some quantities
characterising the state of the system. Their equilibrium values are deter
mined by the fact that, in statistical equilibrium, the entropy S of the whole
system must be a maximum, i.e. we must have X a = for all a, where X a
denotes the derivative
X a = dSldx a . (58.1)
We assume that the system is in a state near to equilibrium. This means that
t The fluxes q and i are independent of the pressure gradient (for given grad fx and grad T), for
the same reason as that given with regard to q in §49.
X See Statistical Physics, §119, Pergamon Press, London 1958.
§58 Coefficients of mass transfer and thermal diffusion 223
all the x a are very little different from their equilibrium values, and the
X a are small. Processes will occur in the system which tend to bring it into
equilibrium. The quantities x a are functions of time, and their rate of change
is given by the time derivatives x a ; we express the latter as functions of X a ,
and expand these functions in series. As far as terms of the first order we have
x a =  ^2yabX b . (58.2)
b
The symmetry principle for the kinetic coefficients states that the y ab (called
the kinetic coefficients) are symmetrical with respect to the suffixes a and b :
Yab = 7ba (58.3)
The rate of change of the entropy S is
S= T,X a x a . (58.4)
Now let the x a themselves be different at different points of the system, i.e.
each volume element have its own values of the x a . That is, we suppose the x a
to be functions of the coordinates. Then, in the expression for S, besides
summing over a we must integrate over the volume of the system:
S =  (^X a x a dV. (58.4a)
J a
It is usually true that the values of the x a at any given point depend only on
the values of the X a at that point. In this case we can write down the
relation between x a and X a for each point in the system, and obtain the same
formulae as previously.f
In the problem under consideration we take as the x a the components of
the vectors i and q— [A. Then we see from a comparison of (57.7) and (58.4a)
that the X a are respectively the components of the vectors (l/T)grad/j
and (1/T 2 ) grad T. The kinetic coefficients y a b are the coefficients of these
vectors in the equations
i^(!^)/^J=JI).
q ,i=ar(^) y r*(i^).
By the symmetry of the kinetic coefficients, we must have /ST 2 = 8T, or
8 = (3T. This is the required relation.
t Strictly speaking, in order to apply the relations obtained for a discrete set of quantities to a
continuous distribution, we should write the integral (58.4a) as a sum over small but finite
regions AV of the body (cf. §132); then the definition of the coefficients yab also involves AV. In
the present case, however, this procedure is unnecessary, since we use only the symmetry of the
kinetic coefficients, and not their actual values.
224 Diffusion §58
We can therefore write the fluxes i and q as
i = a grad/*£ grad T,
q = _j8r grad/*y grad T+/>ti,
with only three independent coefficients a, /?, y. It is convenient to eliminate
grad ju from the expression for the heat flux, replacing it by i and grad T.
Then we have
i = agrad^jSgradr, (58.6)
q = (^+j8T/a)i k grad T, (58.7)
where
K = yp2T/oL. (58.8)
If the diffusion flux i is zero, we have pure thermal conduction. For this
to be so, T and p must satisfy the equation a grad ju + fi grad T = 0, or
adj^+^dT = 0. The integration of this equation gives a relation of the
form/(c, T) — which does not contain the coordinates explicitly. (The
chemical potential is a function of the pressure, as well as of c and T, but in
equilibrium the pressure is constant.) This relation determines the depen
dence of the concentration on the temperature which must hold if there is
no diffusion flux. Moreover, for i = we have from (58.7)
q = — k grad T,
so that k is just the thermal conductivity.
Let us now change to the usual variables p, T and c. We have
grad/z = (dpi dc) v . T grad c + {d[xj '8T) C>P grad T+ (dfi/ dp) cT gradp.
In the last term we can replace the derivative (dpjdp) Ct T by (dVjdc) Pt T,
where V is the specific volume.f Substituting in (58.6) and (58.7), and putting
p \ dc / T,p
(58.9)
P k T DIT = v.(dpjdT) CtV +p y
k P = p(dVldc) PiT l(dpldc) p T , (58.10)
we obtain
i =  P D[gradc+(k T IT) grad T+(k p /p) grad/>], (58.11)
q = [k T (Spl8c) p , T  T(8pldT) p , c +p]i k grad T. (58.12)
The coefficient D is called the diffusion coefficient or mass transfer coefficient]
f The equality of these two derivatives follows from the thermodynamic identity
d<f> = sdT+Vdp + pdc,
where <f> is the thermodynamic potential per unit mass;
(dpjdp) c>T = dm dp 8c = (8VI8c) PfT .
§58 Coefficients of mass transfer and thermal diffusion 225
it gives the diffusion flux when only a concentration gradient is present.
The diffusion flux due to the temperature gradient is given by the thermal
diffusion coefficient UtD\ the dimensionless quantity Ay is called the thermal
diffusion ratio.
The last term in (58.11) need be taken into account only when there is a
considerable pressure gradient in the fluid (caused by an external field, say).
The coefficient kpD may be called the barodiffusion coefficient. It should
be noticed that, by formula (58.10), the dimensionless quantity k p is entirely
determined by thermodynamic properties alone.
In a single fluid there is, of course, no diffusion flux. Hence it is clear that
kr and k p must vanish in each of the two limiting cases c = and c = 1.
The condition that the entropy must increase places certain restrictions on
the coefficients in formulae (58.6) and (58.7). Substituting these formulae
in the expression (57.7) for the rate of change of the entropy, we find
8 r f /c(gradT) 2 C i 2 , „ n *„
Hence it is clear that, besides the condition k > which we already know,
we must have also a > 0. Bearing in mind that the derivative (d[Mldc) Pt T
is always positive,f we therefore find that the diffusion coefficient must be
positive : D > 0. The quantities &t and k p , however, may be either positive
or negative.
We shall not pause to write out the lengthy general equations obtained by
substituting the above expressions for i and q in (57.3) and (57.6). We
shall take only the case where there is no significant pressure gradient, while
the concentration and temperature of the fluid vary so little that the coeffi
cients in the expressions (58.11) and (58.12) may be supposed constant,
although they are in general functions of c and T. Furthermore, we shall
suppose that there is no macroscopic motion in the fluid except that which
may be caused by the temperature and concentration gradients. The velocity
of this motion is proportional to the gradients, and the terms in equations
(57.3) and (57.6) which involve the velocity are therefore quantities of the
second order, and may be neglected. The term — i«grad y. in (57.6) is also of
the second order. Thus we have pdcjdt + div i = 0, pTds/dt + div(q— fii) = 0.
Substituting for i and q the expressions (58.11) and (58.12) (without the
term in gradp), and transforming the derivative dsjdt as follows: J
8s I ds \ 8T I 8s \ 8c c p 8T I 8\i \ 8c
Yt~ \~8f) CtV ~8t \~8c ) T ,p^t ~ ~f ~8t ~ \~8T ) Pte 8t
t See Statistical Physics, §95.
j For
{8sj8c) PtT = d^/dcST = (dfj,l8T) p>c .
226 Diffusion §58
we obtain after a simple calculation
dc/dt = D[Ac+(k T /T)ATl (58.14)
dT/dt  (k T /c p )(dfM/dc) PiT 8c/dt = x A 7 1 . (58.15)
This system of linear equations determines the temperature and concentra
tion distributions in the fluid.
There is a particularly important case where the concentration is small.
When the concentration tends to zero, the diffusion coefficient tends to a
finite constant, but the thermal diffusion coefficient tends to zero. Hence
kr is small for small concentrations, and we can neglect the term krA T
in (58.14), which then becomes the diffusion equation
dc/dt = DAc (58.16)
The boundary conditions on the solution of (58.16) are different in
different cases. At the surface of a body insoluble in the fluid the normal
component of the diffusion flux i = — pD grad c must vanish, i.e. we
must have dc/8n = 0. If, however, there is diffusion from a body which
dissolves in the fluid, equilibrium is rapidly established near its surface,
and the concentration in the fluid adjoining the body is the saturation
concentration Co', the diffusion out of this layer takes place more slowly
than the process of solution. The boundary condition at such a surface is
therefore c = cq. Finally, if a solid surface absorbs the diffusing substance
incident on it, the boundary condition is c — 0; an example of such a case is
found in the study of chemical reactions at the surface of a solid.
Since the equations of pure diffusion (58.16) and of thermal conduction
(50.4) are of exactly the same form, we can immediately apply all the formulae
derived in §§51 and 52 to the case of diffusion, simply replacing T by c and
X by D. The boundary condition for a thermally insulating surface corres
ponds to that for an insoluble surface, while a surface maintained at a constant
temperature corresponds to a soluble surface from which diffusion takes place.
In particular, we can write down, by analogy with (51.6), the following
solution of the diffusion equation :
M
C{r) = J^Dif tM ~ r2lm ' (58  17)
This gives the distribution of the solute at any time, if at time t = it is
all concentrated at the origin (M being the total amount of the solute).
PROBLEM
Determine the barodiffusion coefficient for a mixture of two perfect gases.
Solution. We have for the specific volume V = kT^+n^Jp (the notation is that used
in the second footnote to §57), and the chemical potentials aret
j"l = fl{p, T) + kT log[m/(»i + W 2 )],
/*2 = flip, r) + ^riog[w 2 /(wi + « 2 )].
f See Statistical Physics, §92.
§59 Diffusion of particles suspended in a fluid 227
The numbers n x and n 2 are expressed in terms of the concentration of the first component by
n imi = c, n 2 ms — 1 —c. A calculation using formula (58.10) gives
rlc C 1
k p = (m2mi)c(lc)\ + — .
L «*2 mi]
§59. Diffusion of particles suspended in a fluid
Under the influence of the molecular motion in a fluid, particles suspended
in the fluid move in an irregular manner (called the Brownian motion).
Let one such particle be at the origin at the initial instant. Its subsequent
motion may be regarded as a diffusion, in which the concentration is repre
sented by the probability of finding the particle in any particular volume
element. To determine this probability, therefore, we can use the solution
(58.17) of the diffusion equation. The possibility of this procedure is due to
the fact that, for diffusion in weak solutions (i.e. when c <4 1, which is when
the diffusion equation can be used in the form (58.16)), the particles of the
solute hardly affect one another, and so the motion of each particle can be
considered independently.
Let w(r, t)6r be the probability of finding the particle at a distance between
r and r + dr from the origin at time t. Putting in (58.17) Mjp = 1 and
multiplying by the volume 47rr 2 dr of the spherical shell, we find
w ( r > W r = ^tWn ex P( " r ^ Dt) r2 dr ' (59>1)
Let us determine the mean square distance from the origin at time t.
We have
00
r 2 = j r ^w(r,t)dr. (59.2)
o
The result, using (59.1), is
^ = 6Dt. (59.3)
Thus the mean distance travelled by the particle during any time is propor
tional to the square root of the time.
The diffusion coefficient for particles suspended in a fluid can be cal
culated from what is called their mobility. Let us suppose that some constant
external force f (the force of gravity, for example) acts on the particles. In a
steady state, the force acting on each particle must be balanced by the drag
force exerted by the fluid on a moving particle. When the velocity is small,
the drag force is proportional to it and is v[b, say, where b is a constant.
Equating this to the external force f, we have
v = bt, (59.4)
i.e. the velocity acquired by the particle under the action of the external force
228 Diffusion §59
is proportional to that force. The constant b is called the mobility, and can
in principle be calculated from the equations of fluid dynamics. For example,
for spherical particles of radius R, the drag force is 6ttt)Rv (see (20.14)),
and therefore the mobility is
b = 1I6tt7]R. (59.5)
For nonspherical particles, the drag depends on the direction of motion;
it can be written in the form aacVjc, where aw is a symmetrical tensor (see
(20.15)). To calculate the mobility we have to average over all orientations
of the particle; if a\, az, a% are the principal values of the symmetrical tensor
ciik, then we have
b =  — + — + —. (59.6)
3\ai a 2 as/
The mobility b is simply related to the diffusion coefficient D. To derive
this relation, we write down the diffusion flux i, which contains the usual
term — pD grad c due to the concentration gradient (we suppose the tem
perature constant), and also a term involving the velocity acquired by the
particle owing to the external forces. This latter term is evidently pcv.
Thus
i =  P D grad c + P cM y (59.7)
where we have used the expression (59.4). In a state of thermodynamic
equilibrium, there is no diffusion, and the flux i must be zero. The equili
brium distribution of the concentration of particles suspended in a fluid,
in an external field, is determined by Boltzmann's formula, according to
which c = constant xr p/w , U being the potential energy of the particle
in the external field. Since f = — grad U, we find the equilibrium concen
tration gradient to be grad c = cf/kT. Substituting this in (59.7) and equat
ing i to zero, we have
D = kTb. (59.8)
This is Einstein's relation between the diffusion coefficient and the mobility.
Substituting (59.5) in (59.8), we find the following expression for the
diffusion coefficient for spherical particles:
D = kT/67T7]R. (59.9)
Besides the translatory Brownian motion and diffusion of suspended par
ticles, we may consider also their rotary Brownian motion and diffusion. Just
as the translatory diffusion coefficient is calculated in terms of the drag
force, so the rotary diffusion coefficient can be expressed in terms of the
forces on a particle executing a rotary movement in the fluid.f
t If (nonspherical) particles are suspended in a planeparallel stream with a transverse velocity
gradient, a definite distribution of the particles as regards their orientation in space is established as
a result of the simultaneous action of the orienting forces of fluid dynamics and the disorienting
Brownian motion. For the solution of this problem for ellipsoidal particles, see A. Peterlin and H. A.
Stuart, Zeitschrift fur Physik 112, 1, 1939.
§59 Diffusion of particles suspended in a fluid 229
PROBLEMS
Problem 1. Particles execute Brownian motion in a fluid bounded on one side by a plane
wall; particles incident on the wall "adhere" to it. Determine the probability that a particle
which is at a distance x Q from the wall at time t = will have "adhered" to it after a time t.
Solution. The probability distribution zo(x, t) (where x is the distance from the wall)
is determined by the diffusion equation, with the boundary condition w = for x =
and the initial condition w = §(x—x ) for t = 0. Such a solution is given by formula (52.4)
when T is replaced by vi, x by D, and T (x') in the integrand by S(*'— x ). We then obtain
«<*>') = TTT^ zrT{^p[(xxofl4Dt]exp[(x+xofl4Dt]}.
2s/\TTDt)
The probability of "adhering" to the wall per unit time is given by the diffusion flux Ddw/Bx
for x = 0, and the required probability W(t) over the time t is
Substituting for to, we find
W{t) = D^[dwldx] x = &t.
W{t) = lerf[* /2vW]
Problem 2. Determine the order of magnitude of the time t during which a particle
suspended in a fluid turns through a large angle about its axis.
Solution. The required time t is that during which a particle in Brownian motion moves
over a distance of the order of its linear dimension a. According to (59.3) we have t r*s a 2 /D,
and by (59.9) D ~ kT/ija. Thus t ~ 7]a s lkT.
CHAPTER VII
SURFACE PHENOMENA
§60. Laplace's formula
In this chapter we shall study the phenomena which occur near the surface
separating two continuous media (in reality, of course, the media are separated
by a narrow transitional layer, but this is so thin that it may be regarded as
a surface). If the surface of separation is curved, the pressures near it in the
two media are different. To determine the pressure difference (called the
surface pressure), we write down the condition that the two media are in
thermodynamic equilibrium together, taking into account the properties of
the surface of separation.
Let the surface of separation undergo an infinitesimal displacement.
At each point of the undisplaced surface we draw the normal. The length of
the segment of the normal lying between the points where it intersects the
displaced and undisplaced surfaces is denoted by S£. Then a volume element
between the two surfaces is S£d/, where d/ is a surface element. Let pi
and/>2 be the pressures in the two media, and let S£ be reckoned positive if
the displacement of the surface is towards medium 2 (say). Then the work
necessary to bring about the above change in volume is
j (Pi+p2)8ldf.
The total work 8R done in displacing the surface is obtained by adding to
this the work connected with the change in area of the surface. This part of
the work is proportional to the change S/in the area of the surface, and is aS/,
where a is called the surfacetension coefficient.^ Thus the total work is
8R =  j (p 1 p 2 )8tdf+ a .8f (60.1)
The condition of thermodynamic equilibrium is, of course, that 8R is zero.
Next, let Ri and R% be the principal radii of curvature at a given point of
the surface; we reckon R\ and R 2 as positive if they are drawn into medium 1.
Then the elements of length d/i and d/ 2 on the surface in its principal sections
receive increments (S£/i?i)d/i and (S£/i? 2 )d/ 2 respectively when the surface
undergoes an infinitesimal displacement; here d/i and d/ 2 are regarded as
f For an airwater interface a =725 erg/cm 2 at 20° C; for air and paraffin a = 24 at 20° C.
The surface tension of liquid metals is very large; for instance, at an airmercury interface a = 547
at 175° C; for air and liquid platinum a = 1820 at 2000° C. The surface tension between liquid
helium and its vapour is very small, a =024 at —270° C.
230
§60 Laplace's formula 231
elements of the circumference of circles with radii JRi and R2. Hence the
surface element d/ = d/id/2 becomes, after the displacement,
d/i(l + 8£/12i)d&(l + 8£/122) « d/id&(l + 8£/22i + 8(7*2),
i.e. it changes by S£d/(l/2?i+l/l?2) Hence we see that the total change in
area of the surface of separation is
s/ =Mrt) d/  (60  2)
Substituting these expressions in (60.1) and equating to zero, we obtain
the equilibrium condition in the form
J«K^(s + i)K°
pi
This condition must hold for every infinitesimal displacement of the surface,
i.e. for all S£. Hence the expression in braces must be identically equal to
zero:
»'(k + k) (60  3)
This is Laplace's formula, which gives the surface pressure. We see that,
if Ri and R2 are positive, pi — />2 > 0. This means that the pressure is greater
in the medium whose surface is convex. If Ri = R 2 — 00, i.e. the surface
of separation is plane, the pressure is the same in either medium, as we
should expect.
Let us apply formula (60.3) to investigate the mechanical equilibrium of
two adjoining media. We assume that no external forces act, either on the
surface of separation or on the media themselves. Using formula (60.3), we
can then write the equation of equilibrium as
1 1
1 = constant. (60.4)
Ri R2
Thus the sum of the curvatures must be a constant over any free surface of
separation. If the whole surface is free, the condition (60.4) means that it
must be spherical (for instance, the surface of a small drop, for which the
effect of gravity may be neglected). If, however, the surface is supported
along some curve (for instance, a film of liquid on a solid frame), its shape is
less simple.
When the condition (60.4) is applied to the equilibrium of thin films
supported on a solid frame, the constant on the right must be zero. For the
sum 1/Ri+ l/i?2 must be the same everywhere on the free surface of the film,
while on opposite sides of the film it must have opposite signs, since, if one
side is convex, the other side is concave, and the radii of curvature are the
same with opposite signs. Hence it follows that the equilibrium condition
232 Surface Phenomena §60
for a thin film is
k + k=° < 60  5 >
Let us now consider the equilibrium condition on the surface of a medium
in a gravitational field. We assume for simplicity that medium 2 is simply
the atmosphere, whose pressure may be regarded as constant over the surface,
and that medium 1 is an incompressible fluid. Then we havep2 = constant,
while pi, the fluid pressure, is by (3.2) pi = constant  pgz, the coordinate
z being measured vertically upwards. Thus the equilibrium condition
becomes
1 1 gpz
— + —  H = constant. (60.6)
Ri R% a
It should be mentioned that, to determine the equilibrium form of the
surface of the fluid in particular cases, it is usually convenient to use the
condition of equilibrium, not in the form (60.6), but by directly solving the
variational problem of minimising the total free energy. The internal free
energy of an incompressible fluid depends only on the volume of the fluid, and
not on the shape of its surface. The latter affects, firstly, the surface free
energy J a d/ and, secondly, the energy in the external field (gravity), which
* s gP J z dV. Thus the equilibrium condition can be written
a J df+gp J zdV = minimum. (60.7)
The minimum is to be determined subject to the condition
f dV = constant, (60.8)
which expresses the fact that the volume of the fluid is constant.
The constants a, p and g appear in the equilibrium conditions (60.6)
and (60.7) only in the form cc/gp. This ratio has the dimensions cm 2 . The
length
a = VV*lgp) (60.9)
is called the capillary constant for the substance concerned.f The shape of
the fluid surface is determined by this quantity alone. If the capillary
constant is large compared with the dimension of the medium, we may
neglect gravity in determining the shape of the surface.
In order to find the shape of the surface from the condition (60.4) or
(60.6), we need formulae which determine the radii of curvature, given the
shape of the surface. These formulae are obtained in differential geometry,
t For water (e.g.), a = 0122 cm at 20° C.
§60
Laplace's formula
233
but in the general case they are somewhat complicated. They are consider
ably simplified when the surface deviates only slightly from a plane. We shall
derive the appropriate formula directly, without using the general results of
differential geometry.
Let z = £(#, y) be the equation of the surface ; we suppose that £ is every
where small, i.e. that the surface deviates only slightly from the plane z = 0.
As is well known, the area / of the surface is given by the integral
or, for small £, approximately by
The variation bf is
'JK(2H(
8y
dxdy.
(60.10)
)dx ay.
8y dy )
8% 8%
+
\8ldxdy.
(60.11)
8x 8x
Integrating by parts, we find
s '=J"(2
Comparing this with (60.2), we obtain
1 1 _ / 8^ 8%
#i + jfo~ ~\~8x~^~8f
This is the required formula; it determines the sum of the curvatures of a
slightly curved surface.
When three adjoining media are in equilibrium, the surfaces of separation
are such that the resultant of the surfacetension forces is zero on the common
line of intersection. This condition implies that the surfaces of separation
must intersect at angles (called angles of contact) determined by the values of
the surfacetension coefficients.f
Finally, let us consider the question of the boundary conditions that must
be satisfied at the boundary between two fluids in motion, when the surface
tension forces are taken into account. If the latter forces are neglected, we
have at the boundary between the fluids flj^a^.i* — vi,ik) = 0, which expresses
the equality of the forces of viscous friction on the surface of each fluid.
When the surface tension is included, we have to add on the righthand
side a force determined in magnitude by Laplace's formula and directed
along the normal :
Wfto^ift— nicotic
( X l
= a — + —
\Ri R 2
(60.12)
t See, for instance, Statistical Physics, §145, Pergamon Press, London 1958.
234 Surface Phenomena §60
This equation can also be written
(Pip2)tii = (a' 1>ik  ff' 2>tt )% + a —  + — )m. (60.13)
If the two fluids are both ideal, the viscous stresses a' tk are zero, and we return
to the simple equation (60.3).
The condition (60.13), however, is still not completely general. The reason
is that the surfacetension coefficient a may not be constant over the surface
(for example, on account of a variation in temperature). Then, besides the
normal force (which is zero for a plane surface), there is another force
tangential to the surface. Just as there is a volume force grad/> per unit
volume (see §2) in cases where the pressure is not uniform, so we have here a
tangential force f, = grad a per unit area of the surface of separation. In
this case we take the positive gradient, because the surfacetension forces
tend to reduce the area of the surface, whereas the pressure forces tend to
increase the volume. Adding this force to the righthand side of equation
(60.13), we obtain the boundary condition
T / l 1 \1 doc
/>iZ>2a^— + — J^m = (ff'i iir nttK+; (60.14)
the unit normal vector n is directed into medium 1. We notice that this
condition can be satisfied only for a viscous fluid: in an ideal fluid, a' iJc =
and the lefthand side of equation (60.14) is a vector along the normal,
while the righthand side is in this case a tangential vector. This equality
cannot hold, except of course in the trivial case where both sides are zero.
PROBLEMS
Problem 1. Determine the shape of a film of liquid supported on two circular frames
with their centres on a line perpendicular to their planes, which are parallel; Fig. 31 shows a
crosssection of the film.
Solution. The problem amounts to that of finding the surface having the smallest area
that can be formed by the revolution about the line r = of a curve r — r(z) which passes
between two given points A and B. The area of a surface of rotation is
7m
/dr\2
>>NHl)Y
§60
Laplace's formula
235
It is well known that the minimum of an integral of the form
J L(x, x)
dt
is given by the equation L — x dL/dx = constant. In the present case this leads to
r = Cl V[l + (drldzf],
whence we have by integration r = c x cosh[(«— c 2 )/cj. Thus the required surface (called a
catenoid) is that formed by the revolution of a catenary. The constants c x and c 2 must be
chosen so that the curve r{z) passes through the given points A and B. The value of c 2
depends only on the choice of the origin of z. For the constant c lf however, two values are
obtained, of which the larger must be chosen (the smaller does not give a minimum of the
integral).
When the distance h between the frames increases, it reaches a value for which the equation
for the constant c x no longer has a real root. For greater values of h, only the shape consisting
of one film on each frame is stable. For example, for two frames of equal radius R the catenoid
form is impossible for a distance h between the frames greater than 1 33.R.
Problem 2. Determine the shape of the surface of a fluid in a gravitational field and
bounded on one side by a vertical plane wall. The angle of contact between the fluid and the
wall is 6 (Fig. 32).
z
l
Fig. 32
Solution. We take the coordinate axes as shown in Fig. 32. The plane x = is the plane
of the wall, and z = is the plane of the fluid surface far from the wall. The radii of curvature
of the surface z = z(x) are R x = oo, R 2 = (1 +z' 2 ) i Jz", so that equation (60.6) becomes
2z
a* (1 + *' 2 )*
= constant,
(1)
where a is the capillary constant. For * = oo we must have z = 0, 1/R 2 = 0, and the constant
is therefore zero. A first integral of the resulting equation is
1
= A— .
(2)
V(l + *' 2 ) a'
From the condition at infinity (z = z' = for * = oo) we have A = 1. A second integration
gives
a . V2«
+
•y(»S)
+ Xq.
The constant * must be chosen so that, at the surface of the wall (x = 0), we have
z' — —cot 9 or, by (2), z = h, where h = ay/{\— sin 0) is the height to which the fluid
rises at the wall itself.
236
Surface Phenomena
§60
Problem 3. Determine the shape of the surface of a fluid rising between two parallel
vertical flat plates (Fig. 33).
Fig. 33
Solution. We take the ysplane halfway between the two plates, and the ayplane to
coincide with the fluid surface far from the plates. In equation (1) of Problem 2, which gives
the condition of equilibrium and is therefore valid everywhere on the surface of the fluid
(both between the plates and elsewhere), the conditions at x = oo again give the constant as
zero. In the integral (2), the constant A is now different according as \x\ > id or Ixl < id
(the function *(*) having a discontinuity for x = id). For the space between the plates,
the conditions are z = for x = and *' = cot B for x = id, where $ is the angle of contact.
According to (2) we have for the heights z = *(0) and x x = z(id): z = ax/(Al)
Zi = aV(A—sm 0). Integrating (2), we obtain
So
(Az 2 /a*)dz
V[l(4# 2 /« 2 ) 2 ]
aV(Acos gy=z
= \a
\
cos i dg
V(Acos£y
where ^is a new variable related to z by z = a y/(A cos i). This is an elliptic integral, and
cannot be expressed in terms of elementary functions. The constant A is found from the
condition that z — z t for x = \d, or
in6
..;
cos f d£
V(^cos£)'
The formulae obtained above give the shape of the fluid surface in the space between the
plates. As d *• 0, A tends to infinity. Hence we have for d <^ a
VA
tn—u
cos£d£ = — — cos0,
J \/A
or A = (a 2 /d 2 ) cos 2 0. The height to which the fluid rises is z X z t X (a 2 /d) cos 9; this
formula can also be obtained directly, of course.
Problem 4. A thin nonuniformly heated layer of fluid rests on a horizontal plane solid
surface; its temperature is a given function of the coordinate x in the plane, and (because
the layer is thin) may be supposed independent of the coordinate z across the layer. The
nonuniform heating results in the occurrence of a steady flow, and its thickness £ con
sequently varies in the xdirection. Determine the function C(x).
Solution. The fluid density p and the surface tension a are, together with the temperature
known functions of x. The fluid pressure p = Po + pgtfz), where p is the atmospheric
§61 Capillary waves 237
pressure (the pressure on the free surface) ; the variation of pressure due to the curvature of the
surface may be neglected. The fluid velocity in the layer may be supposed everywhere parallel
to the »axis. The equation of motion is
rjd^/dz 2 = dpjdx = g[d(p£)ldxzd P ldx]. (1)
On the solid surface (z = 0) we have v = 0, while on the free surface (z = £) the boundary
condition (60.14) must be fulfilled; in this case it is rj[dv/dz] z <=z = da/dx. Integrating equa
tion (1) with these conditions, we obtain
ijv = gz(i;lz)d(pi;)ldxigz(3PzZ)dpldxzdoLldx. (2)
Since the flow is steady, the total mass flux through a crosssection of the layer must be
zero :
jvdz = 0.
Substituting (2), we find
dt 2 dp 1 da
In + ££2_L t
dx dx g dx
3A
which determines the function £(x). Integrating, we obtain
gt 2 = 3/>*[f/r*da + constant]. (3)
If the temperature (and therefore p and a) varies only slightly, then (3) can be written
P = Wpolp)* + %*«o)lf>g,
where £ is the value of £ at a point where p = p ^d a = a o
§61. Capillary waves
Fluid surfaces tend to assume an equilibrium shape, both under the action
of the force of gravity and under that of surfacetension forces. In studying
waves on the surface of a fluid in §§12 and 13, we did not take the latter forces
into account. We shall see below that capillarity has an important effect on
gravity waves of small wavelength.
As in §12, we suppose the amplitude of the oscillations small compared
with the wavelength. For the velocity potential we have as before the equa
tion A^ = 0. The condition at the surface of the fluid is now different,
however: the pressure difference between the two sides of the surface is
not zero, as we supposed in §12, but is given by Laplace's formula (60.3).
We denote by I the z coordinate of a point on the surface. Since £ is
small, we can use the expression (60.11), and write Laplace's formula as
\ dx 2 By 2
Here p is the pressure in the fluid near the surface, and po is the constant
external pressure. For p we substitute, according to (12.2),
P = pglpm^t,
238 Surface Phenomena §61
obtaining
Pgt+p—  a + — =0;
dt \dx* dyZJ
for the same reasons as in §12, we can omit the constant p if we redefine <f>.
Differentiating this relation with respect to t, and replacing dt,\dt by tyjdz,
we obtain the boundary condition on the potential <f>:
d<f> 8U d / 8U d 2 4>\
'^ + ^"&(^ + vH for * =0  (6U)
Let us consider a plane wave propagated in the direction of the #axis.
As in §12, we obtain a solution in the form <f> = Ae kz cos(kxcot). The
relation between k and o> is now obtained from the boundary condition
(61.1), and is
co 2 =gk + xkZjp. (61.2)
We see that, for long wavelengths such that k ^ V{gpl<*), or k ^ \\a
(where a is the capillary constant), the effect of capillarity may be neglected,
and we have a pure gravity wave. In the opposite case of short wavelengths,
the effect of gravity may be neglected. Then
o>2 = ajfi/p, ( 61 3 )
Such waves are called capillary waves or ripples. Intermediate cases are
referred to as capillary gravity waves.
Let us also determine the nature of the oscillations of a spherical drop of
incompressible fluid under the action of capillary forces. The oscillations
cause the surface of the drop to deviate from the spherical form. As usual, we
shall suppose the amplitude of the oscillations to be small.
We begin by determining the value of the sum l/i?i+ 1/R 2 for a surface
slightly different from that of a sphere. Here we proceed as in the derivation
of formula (60.11). The area of a surface given in spherical coordinatesf
r, 6, <f> by a function r = r{6, <f>) is
T
A spherical surface is given by r = constant = R (where R is the radius
of the sphere), and a neighbouring surface by r = R+£, where £ is small.
Substituting in (61.4), we obtain
'■'m^m^Q'
sin0d0d^.
t In the remainder of this section </> denotes the azimuthal angle, and we denote the velocity
potential by tfi.
§61 Capillary waves 239
Let us find the variation 6/ in the area when £ changes. We have
2 f H dl dhl 1 dl Bht, )
bf = \{2(R + mt + — —  + —  sin0d0d<£.
J J J I K } 36 dd sin20 d<f> d<f> I Y
Integrating the second term by parts with respect to 0, and the third by
parts with respect to </>, we obtain
rVr 13/ an 1 a 2 n
8f= 2CR + £) s in0— ^— J3£sin0d0d<£
J J J\ K ' sin0 dd\ 86 sin20 ^2 J
If we divide the expression in braces by R(R + 2Q, the resulting coefficient
of S£S/ « 8£R(R+ 2£) sin dddd<f> in the integrand is, by formula (60.2),
just the required sum of the curvatures, correct to terms of the first order in £.
Thus we find
± + ± = !_?£_l(_if£ + _L!(si„^)l. (6i.5)
R x R 2 R R 2 # 2 lsin20 d<p sin0 dd \ ddl)
The first term corresponds to a spherical surface, for which JRi = R% = R.
The velocity potential if/ satisfies Laplace's equation A^ = 0, with a
boundary condition atr = R like that for a plane surface:
dip (2 21 l r l a / an l a^n
pJL + J sin0— + — ■ — \\+po = 0.
P a* 1/2 /?2 /?2Lsin0a0\ 50/ sin20 a^2jj
The constant po + 2«.jR can again be omitted; differentiating with respect to
time and putting dlfdt = v r = a«/»/ar, we have finally the boundary condition
on^:
P
dp RH dr arLsin0a0\ 30 J sin 2 d<f>*]\
for r = R. (61.6)
We shall seek a solution in the form of a stationary wave: tp = er+tflr, 0, ^),
where the function /satisfies Laplace's equation, A/ = 0. As is well known,
any solution of Laplace's equation can be represented as a linear combination
of what are called volume spherical harmonic functions r l Y lm (9, <f>), where
Y lm (d,<f>) are Laplace's spherical harmonics: Y lm (d, <f>) = P, w (cos 6)e im< t>.
Here P, m (cos 0) = sin m d m Pi (cos 0)/d (cos 0) m is what is called an associated
Legendre function, Pi (cos 0) being the Legendre polynomial of order /. As
is well known, / takes all integral values from zero upwards, while m takes the
values 0, ±1, ±2, ..., ±1
Accordingly, we seek a particular solution of the problem in the form
tf, = Ae i(0t r l Pi m (cos dy™*. (61.7)
240 Surface Phenomena §61
The frequency co must be such as to satisfy the boundary condition (61.6).
Substituting the expression (61.7) and using the fact that the spherical har
monics Y lm satisfy
1 3 / . J Y im\ 1 8*Yi m
sin0 dd\ 86/ sin 2 302
we find (cancelling ift)
pa> + fo[2l(l+l)]IR? = 0,
or
o? = o/(/l)(/+2)/p/28. (61.8)
This formula gives the eigenfrequencies of capillary oscillations of a
spherical drop. We see that it depends only on /, and not on m. To a given /,
however, there correspond 21+ 1 different functions (61.7). Thus each
of the frequencies (61.8) corresponds to 21+ 1 different oscillations. Inde
pendent oscillations having the same frequency are said to be degenerate;
in this case we have (21+ l)fold degeneracy.
The expression (61.8) vanishes for / = and / = 1. The value / =
would correspond to radial oscillations, i.e. to spherically symmetrical pulsa
tions of the drop; in an incompressible fluid such oscillations are clearly
impossible. For / = 1 the motion is simply a translatory motion of the drop
as a whole. The smallest possible frequency of oscillations of the drop cor
responds to / = 2, and is
w m in = V(8a//># 3 ). (61.9)
A peculiar wave motion due to surface tension is observed when a thin
layer of viscous fluid flows down a vertical wall. P. L. Kapitza has shown that
these waves must be due to an instability of the original flow that sets in at
comparatively small Reynolds numbers.f
PROBLEMS
Problem 1. Determine the frequency as a function of the wave number for capillary
gravity waves on the surface of a fluid of depth h.
Solution. Substituting in the condition (61.1) <j> = A cos(kxo)t) cosh k(z+h) (cf. §12,
Problem 1), we obtain co 2 = (gk+a.k 3 Jp) tanh kh. For tt>lwe return to formula (61.2)',
while for long waves (kh <^ 1) we have to 2 = ghk 2 + a.hk i Jp.
Problem 2. Determine the damping coefficient for capillary waves.
Solution. Substituting (61.3) in (25.5), we find y = 2 v k 2 /p = 2i?w 4 / 3 /p 1/3 a 2/s .
Problem 3. Find the condition for the stability of a horizontal tangential discontinuity in a
gravitational field, taking account of surface tension (the fluids on the two sides of the sur
face of discontinuity being supposed different).
Solution. Let U be the velocity of the upper fluid relative to the lower. On the original
flow we superpose a perturbation periodic in the horizontal direction, and seek the velocity
t See P. L. Kapitza, Zhurnal eksperimental'nol i teoreticheskoi fiziki 18, 3, 1948.
§62 The effect of adsorbed films on the motion of a liquid 241
potential in the form
<f> = Ae kz COs(&tf — 0)t) in the lower fluid,
<£' = A'e~ kz COS(&V— COt)+ Ux in the upper fluid.
For the lower fluid we have on the surface of discontinuity v z = d<j>Jdz = d£/d£, where t, is a
vertical coordinate in the surface of discontinuity, and for the upper fluid
v ' z = dtffdz = Udydx+dt/dt.
The condition of equal pressures in the two fluids at the surface of discontinuity is
P d4>ldt+pgl*Ptld& = P 'd<f>'ldt+ P 'gt:+ip\v' 2 U 2 );
only terms of the first order in A' need be retained in expanding the expression v'* — U 2 .
We seek the displacement £ in the form t, = asin(kx — cat). Substituting <f>, <f>' and £ in
the above three conditions for z — 0, we obtain three equations from which a, A and A'
can be eliminated, leaving
tyU l\kg{p P ') tfpp'UZ a&
CO =
l + I\ kg(pP) Wpp'U*  a#» I
/ ± VL p+p {p+pf p+p y
p+p
In order that this expression should be real for all k, it is necessary that
V A ^ 4* g (pp')(p+p'ripy 2 .
If this condition does not hold, there are complex a> with a positive imaginary part, and the
motion is unstable.
§62. The effect of adsorbed films on the motion of a liquid
The presence on the surface of a liquid of a film of adsorbed material may
have a considerable effect on the hydrodynamical properties of the surface.
The reason is that, when the shape of the surface changes with the motion of
the liquid, the film is stretched or compressed, i.e. the surface concentration
of the adsorbed substance is changed. These changes result in the appearance
of additional forces which have to be taken into account in the boundary
conditions at the free surface.
Here we shall consider only adsorbed films of substances which may be
regarded as insoluble in the liquid. This means that the substance is entirely
on the surface, and does not penetrate into the liquid. If the adsorbed
substance is appreciably soluble, it is necessary to take into account the
it diffusion of between the surface film and the volume of the liquid when the
concentration of the film varies.
When the adsorbed material is present, the surfacetension coefficient a
is a function of the surface concentration of the material (the amount of it
per unit surface area), which we denote by y. If y varies over the surface,
then the coefficient a is also a function of the coordinates in the surface.
The boundary condition at the surface of the liquid therefore includes
a tangential force, which we have already discussed at the end of §60 (equation
(60.14)). In the present case, the gradient of a can be expressed in terms of
the surface concentration gradient, so that the tangential force on the surface is
ft = (3a/dy)grady. (62.1)
242 Surface Phenomena §62
It has been mentioned in §60 that the boundary condition (60.14), in which
this force is taken into account, can be satisfied only for a viscous fluid.
Hence it follows that, in cases where the viscosity of the liquid is small, and
unimportant as regards the phenomenon under consideration, the presence
of the film can be ignored.
To determine the motion of a liquid covered by a film we must add to the
equations of motion, with the boundary condition (60.14), a further equation,
since we now have another unknown quantity, the surface concentration y.
This further equation is an "equation of continuity", expressing the fact that
the total amount of adsorbed material in the film is unchanged. The actual
form of the equation depends on the shape of the surface. If the latter is
plane, then the equation is evidently
dyldt + 8(yv x )/8x + 8(yv y )/8y = 0, (62.2)
where all quantities have their values at the surface (taken as the ayplane).
The solution of problems of the motion of a liquid covered by an adsorbed
film is considerably simplified in cases where the film may be supposed
incompressible, i.e. we may assume that the area of any surface element of the
film remains constant during the motion.
An example of the important hydrodynamic effects of an adsorbed film is
given by the motion of a gas bubble in a viscous liquid. If there is no film
on the surface of the bubble, the gas inside it moves also, and the drag force
exerted on the bubble by the liquid is not the same as the drag on a solid
sphere of the same radius (see §20, Problem 2). If, however, the bubble is
covered by a film of adsorbed material, it is clear from symmetry that the
film remains at rest when the bubble moves. For a motion in the film could
occur only along meridian lines on the bubble surface, and the result would be
that material would continually accumulate at one of the poles (since the
adsorbed material does not penetrate into the liquid or the gas); this is
impossible. Besides the velocity of the film, the gas velocity at the surface
of the bubble must also be zero, and with this boundary condition the gas
in the bubble must be entirely at rest. Thus a bubble covered by a film moves
like a solid sphere and, in particular, the drag on it (for small Reynolds
numbers) is given by Stokes' formula. This result is due to V. G. Levich,
who also gave the solutions to the following Problems.f
PROBLEMS
Problem 1. Two vessels are joined by a long deep channel of width a and length / with
plane parallel walls. The surface of the liquid in the system is covered by an adsorbed film,
and the surface concentrations y x and y 2 of the film in the two vessels are different. There
results a motion near the surface of the liquid in the channel. Determine the amount of
film material transported by this motion.
Solution. We take the plane of one wall of the channel as the xzplane, and the surface
of the liquid as the #yplane, so that the *axis is along the channel; the liquid is in the region
t For a more detailed account see V. G. Levich, Physicochemical Hydrodynamics (Fiziko
khimicheskaya gidrodinamika), Moscow 1952.
§62 The effect of adsorbed films on the motion of a liquid 243
z < 0. There is no pressure gradient, so that the equation of steady flow is (cf. §17)
d 2 v d 2 v
+ = o, (1)
dy* 3*2
where v is the liquid velocity, which is evidently in the acdirection. There is a concentration
gradient dy/d* along the channel. At the surface of the liquid in the channel we have the
boundary condition
7) dvjdz = doc/dx for z = 0. (2)
At the channel walls the liquid must be at rest, i.e.
V = for y = and y = a. (3)
The channel depth is supposed infinite, and so
V = for Z > — 00. (4)
Particular solutions of equation (1) which satisfy the conditions (3) and (4) are
v n — constant X exp[(2« + l)iTz[a] sin(2»+l)7ry/a,
with n integral. The condition (2) is satisfied by the sum
4a da^ exp[(2«+l)7r#/a] sm{2n + \)iryja
yS^ (2« + l)2 '
The amount of film material transferred per unit time is
f 8a 2 / ^ 1 \ da
the motion being in the direction of a increasing. The value of Q must obviously be constant
along the channel. Hence we can write
da If da l? 1 ,
y — = constant =  y— dx=  yd<x,
dx I J dx I J
«s
where a x = a(y x ), <% == a(y 2 ), and we assume that a x > a^ Thus we have finally
8a 2 /^ 1 \ f a 2 f 1
Q = ( > y da = 027— y da.
T ? /7r3\^(2«+l)3/ J Y r)lj r
Problem 2. Determine the damping coefficient for capillary waves on the surface of a
liquid covered by an adsorbed film.
Solution. If the viscosity of the liquid is not too great, the stretching (tangential) forces
exerted on the film by the liquid are small, and the film may therefore be regarded as in
compressible. Accordingly, we can calculate the energy dissipation as if it took place at a
244 Surface Phenomena §62
solid wall, i.e. from formula (24.14). Writing the velocity potential in the form
eh = J)q gikx—iat g—kz
we obtain for the dissipation per unit area of the surface
T
En.m =  ^{lpt]0))\hj>o\ 2 .
The total energy (also per unit area) is
e = p\^dz = y\k^ik.
The damping coefficient is (using (61.3))
0,7/<y/2 £7/4^1/2^/4
2^2<x.v s p 1/6 ~ 2V2p 3 ' 4 '
The ratio of this quantity to the damping coefficient for capillary waves on a clean surface
(§61, Problem 2) is (a/>/£i? 2 ) l/4 /4V2, and is large compared with unity unless the wavelength
is extremely small. Thus the presence of an adsorbed film on the surface of a liquid leads to
a marked increase in the damping coefficient.
7 =
CHAPTER VIII
SOUND
§63. Sound waves
We proceed now to the study of the flow of compressible fluids, and begin
by investigating small oscillations; an oscillatory motion of small amplitude
in a compressible fluid is called a sound wave. At each point of the fluid,
a sound wave causes alternate compression and rarefaction.
Since the oscillations are small, the velocity v is small also, so that the term
(v«grad)v in Euler's equation may be neglected. For the same reason, the
relative changes in the fluid density and pressure are small. We can write
the variables p and p in the form
p=po+p'> p = po+p, (63.1)
where po and po are the constant equilibrium density and pressure, and p
and/>' are their variations in the sound wave (p <^ po, p' 4 Po). The equation
of continuity dpldt+div(pv) = 0, on substituting (63.1) and neglecting small
quantities of the second order (/>', p' and v being of the first order), becomes
dp'/dt+podivv = 0. (63.2)
Euler's equation
3v/3* + (vgrad)v = (l/p)gradp
reduces, in the same approximation, to
0v/0*+(l/po)gradp' = 0. (63.3)
The condition that the linearised equations of motion (63.2) and (63.3)
should be applicable to the propagation of sound waves is that the velocity
of the fluid particles in the wave should be small compared with the velocity
of sound: v <^ c. This condition can be obtained, for example, from the
requirement that p' <^ po (see formula (63.12) below).
Equations (63.2) and (63.3) contain the unknown functions v, p' and p' .
To eliminate one of these, we notice that a sound wave in an ideal fluid is,
like any other motion in an ideal fluid, adiabatic. Hence the small change
p' in the pressure is related to the small change p' in the density by
p' = (dp!d P o) s p f . (63.4)
Replacing p according to this equation in (63.2), we find
dp'l8t+po(dpldp ) s divv = 0. (63.5)
The two equations (63.3) and (63.5), with the unknowns v and p', give a
complete description of the sound wave.
9 245
246 Sound §63
In order to express all the unknowns in terms of one of them, it is con
venient to introduce the velocity potential by putting v = grad <f>. We
have from equation (63.3)
p' = ptyjdt, (63.6)
which relates p' and (f> (here, and henceforward, we omit for brevity the
suffix inpo and po). We then obtain from (63.5) the equation
dmdt2c*A<f> = 0, (63.7)
which the potential <f> must satisfy; here we have introduced the notation
c = Vilify)*. (63.8)
An equation of the form (63.7) is called a wave equation. Applying the
gradient operator to (63.7), we find that each of the three components of the
velocity v satisfies an equation of the same form, and on differentiating
(63.7) with respect to time we see that the pressure p' (and therefore p)
also satisfies the wave equation.
Let us consider a sound wave in which all quantities depend on only one
coordinate (x, say). That is, the flow is completely homogeneous in the
ysplane. Such a wave is called a. plane wave. The wave equation (63.7)
becomes
8P4>/dx*(llc*)dmdt* = 0. (63.9)
To solve this equation, we replace x and t by the new variables £ = x — ct,
t] = x+ct. It is easy to see that in these variables (63.9) becomes
d 2 (f>ldir]dg = 0. Integrating this equation with respect to £, we find
8<f>jdr) = F(rf), where F(rj) is an arbitrary function of 17. Integrating again,
we obtain <f> = /i(£) +/2(i?)i where /1 and fa are arbitrary functions of their
arguments. Thus
<f> =fi(xct)+f 2 (x+ct). (63.10)
The distribution of the other quantities (p', />', v) in a plane wave is given by
functions of the same form.
For definiteness, we shall discuss the density, p = fi(x—ct)+f2(x+ct).
For example, let/2 = 0, so that p = fi(x—cf). The meaning of this solution
is evident: in any plane x = constant the density varies with time, and at any
given time it is different for different x, but it is the same for coordinates x
and times t such that x—ct= constant, or x = constant + ct. This means
that, if at some instant t — and at some point the fluid density has a certain
value, then after a time t the same value of the density is found at a distance
ct along the #axis from the original point. The same is true of all the other
quantities in the wave. Thus the pattern of motion is propagated through
the medium in the ^direction with a velocity c; c is called the velocity of
sound.
Thus/i(«— ct) represents what is called a travelling plane wave propagated
in the positive direction of the #axis. It is evident that/^+c*) represents
a wave propagated in the opposite direction.
§63 Sound waves 247
Of the three components of the velocity v = grad in a plane wave, only
v x — d^Jdx is not zero. Thus the fluid velocity in a sound wave is in the
direction of propagation. For this reason sound waves in a fluid are said to
be longitudinal.
In a travelling plane wave, the velocity v x — v is related to the pressure p'
and the density p in a simple manner. Putting <f> = f(x—ct), we find
v — d<f>Jdx = f'(x—ci) and p' — — pd<j>jdt = pcf'{x—ct). Comparing the
two expressions, we find
v = p'/pc. (63.11)
Substituting here from (63.4) p' = c 2 p, we find the relation between the
velocity and the density variation :
v = cp'/p. (63.12)
We may mention also the relation between the velocity and the temperature
oscillations in a sound wave. We have 7" = {dT\dp\p' and, using the well
known thermodynamic formula (dT[dp) s = (Tlcp)(dV/dT)p and formula
(63.11), we obtain
T = tfTv/cp, (63.13)
where /? = {\jV){dVjdT) p is the coefficient of thermal expansion.
Formula (63.8) gives the velocity of sound in terms of the adiabatic
compressibility of the fluid. This is related to the isothermal compressibility
by the thermodynamic formula
(dp/dp), = (c P lc v )(8pldp) T . (63.14)
Let us calculate the velocity of sound in a perfect gas. The equation of state
ispV = pip = RT/p., where R is the gas constant and p. the molecular weight.
We obtain for the velocity of sound the expression
c = V(yRT/p,), (63.15)
where y denotes the ratio Cp\c v .\ Since y usually depends only slightly on the
temperature, the velocity of sound in the gas may be supposed proportional
to the square root of the temperature. For a given temperature it does not
depend on the pressure.
What are called monochromatic waves are a very important case. Here all
quantities are just periodic (harmonic) functions of the time. It is usually
convenient to write such functions as the real part of a complex quantity (see
the beginning of §24). For example, we put for the velocity potential
<f> = re[<f>o(x,y,z)e^l (63.16)
where w is the frequency of the wave. The function <£o satisfies the equation
Ah + Mc^o = 0, (63.17)
which is obtained by substituting (63.16) in (63.7).
t It is useful to note that the velocity of sound in a gas is of the same order of magnitude as the
mean thermal velocity of the molecules.
248 Sound §63
Let us consider a monochromatic travelling plane wave, propagated in the
positive direction of the #axis. In such a wave, all quantities are functions
oi x—ct only, and so the potential is of the form
cf> = re{^4 exp[ico(tx/c)]} y (63.18)
where A is a constant called the complex amplitude. Writing this as A = ae i<x
with real constants a and a, we have
(f> = acos(coxlc—cot + <x). (63.19)
The constant a is called the amplitude of the wave, and the argument of the
cosine is called the phase. We denote by n a unit vector in the direction of
propagation. The vector
k = (o)/c)n = (27r/A)n (63.20)
is called the wave vector. In terms of this vector (63.18) can be written
<j> = re{A exp[)'(k.r cot)]}. (63.21)
Monochromatic waves are very important, because any wave whatsoever
can be represented as a sum of superposed monochromatic plane waves
with various wave vectors and frequencies. This decomposition of a wave
into monochromatic waves is simply an expansion as a Fourier series or inte
gral (called also spectral resolution). The terms of this expansion are called
the monochromatic components or Fourier components of the wave.
PROBLEMS
Problem 1 . Determine the velocity of sound in a nearly homogeneous twophase system
consisting of a vapour with small liquid droplets suspended in it (a "wet vapour"), or a liquid
with small vapour bubbles in it. The wavelength of the sound is supposed large compared
with the size of the inhomogeneities in the system.
Solution. In a twophase system, p and T are not independent variables, but are related
by the equation of equilibrium of the phases. A compression or rarefaction of the system is
accompanied by a change from one phase to the other. Let x be the fraction (by mass) of
phase 2 in the system. We have
S = (lx)Si + XS2y
V = (lx)Vi + xV 2 , (
where the suffixes 1 and 2 distinguish quantities pertaining to the pure phases 1 and 2. To
calculate the derivative (dV/dp)„ we transform it from the variables p, s to p, x, obtaining
(8V/8p), = (dV/8p)x—(dVjdx) ]> (dsldp)xK8sldx) ]> . The substitution (1) then gives
l!L\ = J^_^zZl^1 +(1 _4^l_^zZl^1. (2)
\ dp I s L dp s 2 si dp J L dp s 2 si dp J
The velocity of sound is obtained from (1) and (2), using formula (63.8).
Expanding the total derivatives with respect to the pressure, introducing the latent heat
of the transition from phase 1 to phase 2 (q = T(s 2 — s x )), and using the ClapeyronClausius
equation for the derivative dp/dT along the curve of equilibrium (dp/dT = p/T(V 2 — Vi)),
we obtain the expression in the first brackets in (2) in the form
/ 8V 2 \ 2T I dV 2 \ Tcj?
(wi + r(^i {V2  Vl) ^ Vl)K
The second bracket is transformed similarly.
§64 The energy and momentum of sound waves 249
Let phase 1 be the liquid and phase 2 the vapour; we suppose the latter to be a perfect
gas, and neglect the specific volume V t in comparison with V z . If * <^ 1 (a liquid containing
some bubbles of vapour), the velocity of sound is found to be
c = qppVxIRTy/ifaT), (3)
where jR is the gas constant and p the molecular weight. This velocity is in general very
small ; thus, when vapour bubbles form in a liquid {cavitation), the velocity of sound undergoes
a sudden sharp decrease.
If 1 — x <^ 1 (a vapour containing some droplets of liquid), we obtain
1 ix, 2 C V ,T
c 2 RT q q* v '
Comparing this with the velocity of sound in the pure gas (63.15), we find that here also the
addition of a second phase reduces the value of c, though by no means so markedly.
As * increases from to 1 , the velocity of sound increases monotonically from the value (3)
to the value (4). For # = and * = 1 it changes discontinuously as we go from a onephase
system to a twophase system. This has the result that, for values of x very close to zero or
unity, the usual linear theory of sound is no longer applicable, even when the amplitude of
the sound wave is small; the compressions and rarefactions produced by the wave are in
this case accompanied by a change between a onephase and a twophase system, and the
essential assumption of a constant velocity of sound no longer holds good.
Problem 2. Determine the velocity of sound in a gas heated to such a high temperature
that the pressure of equilibrium blackbody radiation becomes comparable with the gas
pressure.
Solution. The pressure is p = nkT+iakT 4 , and the entropy is
s = (klm)\og(T*/n) + akT3ln.
In these expressions the first terms relate to the particles, and the second terms to the radia
tion; n is the number density of particles, m their mass, k Boltzmann's constant, and
a = 4ir 2 k?l45h 3 c 3 .\ The density of matter is not affected by the blackbody radiation, so that
P = mn. The velocity of sound, denoted here by w to distinguish it from that of light, is
d(p,s) d(p yS ) i d( P ,s)
u 2 =
/■
d(p,s) d(n,T)l d(n,T)
where the derivatives have been written in Jacobian form. Evaluating the Jacobians, we have
SkT r 2*276 ,
m 2 = 1+ .
3m L 5n(n + 2aT3)\
§64. The energy and momentum of sound waves
Let us derive an expression for the energy of a sound wave. According to
the general formula, the energy in unit volume of the fluid is pe + ^pv 2 .
We now substitute p = po + p', e = eo + e', where the primed letters denote
the deviations of the respective quantities from their values when the fluid
is at rest. The term %pv 2 is a quantity of the third order. Hence, if we take
only terms up to the second order, we have
poco+p — — + ip 2 — —  + ipov 2 .
opo Opo*
f See, for instance, Statistical Physics, §60, Pergamon Press, London 1958.
250 Sound §64
The derivatives are taken at constant entropy, since the sound wave is adiaba
tic. From the thermodynamic relation de = Tds—pdV = Tds+(p}p 2 )dp we
have [d(p€)fdp] s = c+pjp = w, and the second derivative is
[82(pe)l8p] s = (8w/8p) s = (dw/dp) s (8pldp) s = c*\p.
Thus the energy in unit volume of the fluid is
/>o*o + wop' + hc 2 p 2 /p + Ipov 2 .
The first term (/»oeo) in this expression is the energy in unit volume when
the fluid is at rest, and does not relate to the sound wave. The second term
(toop) is the change in energy due to the change in the mass of fluid in unit
volume. This term disappears in the total energy, which is obtained by
integrating the energy over the whole volume of the fluid: since the total mass
of fluid is unchanged, we have
jpdV=j Po dV, or j p' dV = 0.
Thus the total change in the energy of the fluid caused by the sound wave is
given by the integral
f (ip0V* + fr2p'2/p )dV.
The integrand may be regarded as the density E of sound energy:
E = Jpooa+fcya/po. (64.1)
This expression takes a simpler form for a travelling plane wave. In such
a wave p = po v\c (see (63.12)), and the two terms in (64.1) are equal, so that
E = p v z . (64.2)
In general this relation does not hold. A similar formula can be obtained only
for the (time) average of the total sound energy. It follows immediately from
a wellknown general theorem of mechanics, that the mean total potential
energy of a system executing small oscillations is equal to the mean total
kinetic energy. Since the latter is, in the case considered,
ijpo^dr,
we find that the mean total sound energy is
j EdV = j potfdV. (64.3)
If a nonmonochromatic wave is represented as a series of monochromatic
waves, the mean energy is equal to the sum of the mean energies of the
monochromatic components. For, if v is represented as a sum of terms of
§64 The energy and momentum of sound waves 25 1
various frequencies, v 2 will contain both the square of each term and the
products of terms of different frequencies. These products contain factors of
the form «*<"*»'>', which are periodic functions of time. But the mean value
of a periodic function is zero, and these terms therefore vanish. Thus the
mean energy contains only terms in the mean squares of the monochromatic
components.
Next, let us consider some volume of a fluid in which sound is propagated,
and determine the mean flux of energy through the closed surface bounding
this volume. The energy flux density in the fluid is, by (6.3), pv(£v 2 +w).
In the present case we can neglect the term in v 2 y which is of the third order.
Hence the mean energy flux density in the sound wave is pzov. Substituting
w = wo + w', we have pwv = zoopv+pw'v. For a small change w' in the
heat function we have w' = (dzv/dp) s p'. Since {dwjdp) 8 = 1/p, it follows that
to' = p'jp and pzov = zoopv+p'v. The total energy flux through the surface
in question is
<j> (wopv+'pv)'df.
However, since the total quantity of fluid in the volume considered is un
changed on the average, the time average of the mass flux through the closed
surface must be zero. Hence the energy flux is simply
(U'vdf.
We see that the mean sound energy flux is represented by the vector
q = ~pv. (64.4)
It is easy to verify that the relation
dE/dt + div(p'v) = (64.5)
holds. In this form the equation gives the law of conservation of the sound
energy, with the vector q = p'v taking the part of the sound energy flux.
Thus the expression is valid not only for the mean flux but also for the flux
at any instant.
In a travelling plane wave the pressure variation is related to the velocity
by P' — c PoV' Introducing the unit vector n in the direction of propagation
of the wave (which is the same as the direction of the velocity v), we obtain
q = cpov 2 n, or
q = cEn. (64.6)
Thus the energy flux density in a plane sound wave equals the energy density
multiplied by the velocity of sound, a result which was to be expected.
Let us now consider a sound wave which, at any given instant, occupies a
finite region of spacef (a wave packet), and determine the total momentum of
f Nowhere bounded by solid walls.
252 Sound §64
the fluid in the wave. The momentum of unit volume of fluid is equal to the
mass flux density j = p\. Substituting p = po + p', we have j = pov+p'v.
The density change is related to the pressure change by p = p'jc 2 . Using
(64.4), we therefore obtain
j = pov+q/c*. (64.7)
Since we have potential flow in a sound wave, we can write v = grad <£ ; it
should be emphasised that this result is not a consequence of the approxi
mations made in deriving the linear equations of motion in §63, since a solu
tion such that curl v = is an exact solution of Euler's equations. We
therefore have j = po grad + q/c 2 . The total momentum in the wave equals
the integral J j dV over the volume occupied by the wave. The integral of
grad <f> can be transformed into a surface integral,
j grad<f>dV = j><f>df,
and is zero, since <f> is zero outside the volume occupied by the wave. Thus the
total momentum of the wave is
JjdF = (1/^2) j qdF. (64.8)
This quantity is not, in general, zero. The existence of a nonzero total
momentum means that there is a transfer of matter. We therefore conclude
that the propagation of a soundwave packet is accompanied by the transfer
of fluid. This is a secondorder effect (since q is a secondorder quantity).
Finally, let us calculate the mean value of the pressure change p' in a
sound wave. In the first approximation, corresponding to the usual linearised
equations of motion, p' is a function which periodically changes sign, and the
mean value of />' is zero. This result, however, ceases to hold if we go to
higher approximations. If we take only secondorder quantities, p' can be
expressed in terms of quantities calculated from the linear sound equations, so
that it is not necessary to solve directly the nonlinear equations of motion
obtained when terms of higher order are taken into account.
We start from Bernoulli's equation : to + \v 2 + d<j>jdt = constant, and average
it with respect to time. The mean value of the time derivative d$\dt is zero.f
Putting also to = too + w' and including too in the constant, we obtain
to' + \v 2 = constant. We suppose that the wave is propagated in an infinite
volume of fluid but is damped at infinity, i.e. v, to', etc. are zero at infinity.
t By the general definition of the mean value, we have for the mean derivative of any function /(*)
T
1 fd/
d//d* = lim — dt = hm
•" r*. 2T J dt r»»
f(T)f{T)
IT
T
If the function /(*) remains finite for all t , the limit is zero, so that d//d* = 0.
§65 Reflection and refraction of sound waves 253
Since the constant is the same in all space, it must evidently be zero, so that
^+1^2 = o. (64.9)
We next expand zv' in powers of p', and take only the terms up to the second
order :
to' = {dw\dp) s p' +\{d 2 w\dp 2 ) s p' 2 \
since (ckvjdp) s = 1/p, we have
p' p' 2 / dp \ p' p' 2
w' =
po 2po 2 \8p/ s po 2c 2 po 2
Substituting this in (64.9) gives
y= Ip^+^Jlpoc 2 = ipo^+ P^^/lpo, (64.10)
which determines the required mean value. The expression on the right is a
secondorder quantity, and is calculated by using the />' and v obtained from
the solution of the linearised equations of motion. The mean density is
J' = {dp\dp*) s p' +\{d 2 p\dp* 2 ) s J 2 . (64.11)
If the wave may be regarded as a travelling plane wave in the volume
concerned, then v = cp'fpo, so that v 2 = c 2 p' 2 lpo 2 , and the expression
(64.10) is zero, i.e. the mean pressure variation in a plane wave is an effect
of higher order than the second. The density variation p = %(d 2 pldpo 2 ) a P' 2
is not zero, however. (We may mention that the derivative (d 2 p[dpo 2 ) s is in
fact always negative, and therefore p < in a travelling wave.) In the same
approximation, we have for the mean value of the momentum flux density
tensor in a travelling plane wave p8i]c+ pvivjc = po&ik + ptfViVjc. The first term
is the equilibrium pressure and does not relate to the sound wave. In the
second term, we introduce the unit vector n in the direction of v (the same
as the direction of propagation of the wave), and, using (64.2), obtain for the
momentum flux density in a sound wave
U ik = Entn k . (64.12)
If the wave is propagated in the ^direction, only the component U X x = E
is not zero. Thus, in this approximation, there is in the plane sound wave only
an ^component of the mean momentum flux, and this is transmitted in the
^direction.
§65. Reflection and refraction of sound waves
When a sound wave is incident on the boundary between two different fluid
media, it undergoes reflection and refraction. This means that, in addition to
254 Sound §65
the incident wave, two more appear; one (the reflected wave) is propagated
back into the first medium from the surface of separation, and the other (the
refracted wave) is propagated into the second medium. Consequently, the
motion in the first medium is a combination of two waves (the incident and
the reflected), whereas in the second medium there is only one, the refracted
wave.
The relation between these three waves is determined by the boundary
conditions at the surface of separation, which require the pressures and normal
velocity components to be equal.
Let us consider the reflection and refraction of a monochromatic longitudi
nal wave at a plane surface separating two media, which we take as the yz
plane. It is easy to see that all three waves have the same frequency co and
the same components k v , k z of the wave vector, but not the same component
k x perpendicular to the plane of separation. For, in an infinite homogeneous
medium, a monochromatic wave with constant k and a> satisfies the equations
of motion. The presence of a boundary introduces only some boundary con
ditions, which in the case considered apply at x = 0, i.e. do not depend on
the time or on the coordinates y, z. Hence the dependence of the solution
on t, y and z remains the same in all space and time, i.e. o>, ky, and k z are
the same as in the incident wave.
From this result we can immediately derive the relations which give the
directions of propagation of the reflected and refracted waves. Let the plane
of the incident wave be the ryplane. Then k z = in the incident wave, and
the same must be true of the reflected and refracted waves. Thus the direc
tions of propagation of the three waves are coplanar.
Let 8 be the angle between the direction of propagation of the wave and
the araxis. Then, from the equality of ky = (oi/c) sin for the incident
and reflected waves, it follows that
h = 0i', (65.1)
i.e. the angle of incidence d\ is equal to the angle of reflection #i'. From a
similar equation for the incident and refracted waves it follows that
sin#i/sin02 = £1/^2. (65.2)
which relates the angle of incidence Q\ to the angle of refraction 9% (c\ and c%
being the velocities of sound in the two media).
In order to obtain a quantitative relation between the intensities of the
three waves, we write the respective velocity potentials as
<f>i — Aiexp[ico{(x/ci) cosdi + (yJci) sindi — t}],
<f>i = A\ exp[/ct){( — x\c\) cosdi + {yjci) sin0i — *}],
<f> 2 = ^2exp[/cy{(«r/c2)cos^2 + (j/^2)sin^2— *}]•
On the surface of separation (x = 0) the pressure (p — —pdtf>jdt) and the
§65 Reflection and refraction of sound waves 255
normal velocities (v x = d<f>ldx) in the two media must be equal; these con
ditions lead to the equations
cos 6i M , v cos 62 m
px{A x + A{) = p 2 A 2y {A x  At!) = A 2 .
C\ C 2
The reflection coefficient R is defined as the ratio of the (time) average energy
flux densities in the reflected and incident waves. Since the energy flux
density in a plane wave is cpv 2 , we have R = cipivi 2 lapivi 2 = ^4i' 2 /^i 2 .
A simple calculation gives
\ />2 tan 02 + pi tan 0i /
The angles 0\ and 2 are related by (65.2); expressing 2 in terms of 0i, we
can put the reflection coefficient in the form
f p 2 C 2 COS 0!  piVJC! 2  C 2 2 sin 2 0^ 1 2
[ p 2 c 2 cos 0! + piV( c i 2  c 2 2 sin 2 0i) J
For normal incidence (#i = 0), this formula gives simply
R = (^Zf^)\ (65.5)
\P2C2 + PlCl/
For an angle of incidence such that
pl 2 (Cl 2 C2 2 )
the reflection coefficient is zero, i.e. the wave is totally refracted. This can
happen if c\ > C2 but p 2 c 2 > pici, or if both inequalities are reversed.
PROBLEM
Determine the pressure exerted by a sound wave on the boundary separating two fluids.
Solution. The sum of the total energy fluxes in the reflected and refracted waves must
equal the incident energy flux. Taking the energy flux per unit area of the surface of separa
tion, we can write this condition in the form CxE^cos 0x = c^'cos ^ 1 +tr 2 JE , s cos Zt where
E lt Ei and E t are the energy densities in the three waves. Introducing the reflection coefficient
R = Ei/E u we therefore have
— C\ cos 01 —
C2 COS 02
The required pressure p is determined as the ^component of the momentum lost per unit
time by the sound wave (per unit area of the boundary). Using the expression (64.12) for
the momentum flux density tensor in a sound wave, we find
p = Ei cosWx + Ek' cos 2 0!E 2 cos 2 2 .
Substituting for E 2 , introducing R and using (65.2), we obtain
p = Ex sin 0i cos 0i[(l + R) cot 0i  (1  R) cot 2 ].
256 Sound §66
For normal incidence {6 1 = 0), we find, using (65.5),
^ I" Pl 2 d 2 +P2 2 C2 2  2p 1 p 2 C 1 2 l
L (j>lCl+p2C2) 2 J'
§66. Geometrical acoustics
A plane wave has the distinctive property that its direction of propagation
and its amplitude are the same in all space. An arbitrary sound wave, of
course, does not possess this property. However, cases can occur where a
sound wave that is not plane may still be regarded as plane in any small
region of space. For this to be so it is evidently necessary that the amplitude
and the direction of propagation should vary only slightly over distances of
the order of the wavelength.
If this condition holds, we can introduce the idea of rays, these being lines
such that the tangent to them at any point is in the same direction as the
direction of propagation; and we can say that the sound is propagated along
the rays, and ignore its wave nature. The study of the laws of propagation
of sound in such cases is the task of geometrical acoustics. We may say that
geometrical acoustics corresponds to the limit of small wavelengths, A » 0.
Let us derive the basic equation of geometrical acoustics, which determines
the direction of the rays. We write the wave velocity potential as
<f> = aeW. (66.1)
In the case where the wave is not plane but geometrical acoustics can be
applied, the amplitude a is a slowly varying function of the coordinates and
the time, while the wave phase ift is "almost linear" (we recall that in a plane
wave ifs = k»r— wt+a., with constant k and co). Over small regions of space
and short intervals of time, the phase ip may be expanded in series; up to
terms of the first order we have
tfi = ifjo + rgradip+tdipjdt.
In accordance with the fact that, in any small region of space (and during
short intervals of time), the wave may be regarded as plane, we define the
wave vector and the frequency at each point as,
k = dt/tjdr = grad«/r, co = di/i/dt. (66.2)
The quantity if/ is called the eikonal.
In a sound wave we have w 2 jc 2 = k 2 = k x 2 +k y 2 + k g 2 . Substituting (66.2),
we obtain the basic equation of geometrical acoustics :
If the fluid is not homogeneous, the coefficient 1/c 2 is a function of the co
ordinates.
§66 Geometrical acoustics 257
As we know from mechanics, the motion of material particles can be
determined by means of the HamiltonJacobi equation, which, like (66.3),
is a firstorder partial differential equation. The quantity analogous to ift
is the action S of the particle, and the derivatives of the action determine the
momentum p = dSfdr and the Hamilton's function (the energy) of the particle
H =  dSjdt\ these formulae are similar to (66.2). We know, also, that the
HamiltonJacobi equation is equivalent to Hamilton's equations
p =  dH/dr, v = r = dH/dp.
From the above analogy between the mechanics of a material particle and
geometrical acoustics, we can write down similar equations for rays:
k = dcoldr, i = dco/dk. (66.4)
In a homogeneous isotropic medium co = ck with c constant, so that k = 0,
r = en (n being a unit vector in the direction of k), i.e. the rays are propagated
in straight lines with a constant frequency to, as we should expect.
The frequency, of course, remains constant along a ray in all cases where
the propagation of sound occurs under steady conditions, i.e. the properties of
the medium at each point in space do not vary with time. For the total time
derivative of the frequency, which gives its rate of variation along a ray, is
dcofdt = dcoldt + rdcjjdr+kdcoldk. On substituting (66.4), the last two
terms cancel, and in a steady state dcofdt = 0, so that dcofdt = 0.
In steady propagation of sound in an inhomogeneous medium at rest
co = ck, where c is a given function of the coordinates. The equations
(66.4) give
r = en, k = k grade. (66.5)
The magnitude of the vector k varies along a ray simply according to k = cofc
(with co constant). To determine the change in direction of n we put
k = conic in the second of (66.5): con/e  (con/e 2 )(r . grade) = k grade,
whence dn/d* = grade + n(n« grade). Introducing the element of length
along the ray dl — c dt, we can rewrite this equation
dn/d/ =  (1/e) grade + n(ngrade)/e. (66.6)
This equation determines the form of the rays; n is a unit vector tangential
to a ray.f
If equation (66.3) is solved, and the eikonal iff is a known function of
coordinates and time, we can then find also the distribution of sound inten
sity in space. In steady conditions, it is given by the equation div q =
(q being the sound energy flux density), which must hold in all space except
t As we know from differential geometry, the derivative dn/d/ along the ray is equal to N/R, where
N is a unit vector along the principal normal and R is the radius of curvature of the ray. The expres
sion on the righthand side of (66.6) is, apart from a factor 1/c, the derivative of the velocity of sound
along the principal normal; hence we can write the equation as l/ic = (l/c)N«grad c. The rays
bend towards the region where c is smaller.
258 Sound §66
at sources of sound. Putting q = cEn, where E is the sound energy density
(see (64.6)), and remembering that n is a unit vector in the direction of
k = grad ifj, we obtain the equation
6iv(cE grad 0/grad 0) = 0, (66.7)
which determines the distribution of E in space.
The second formula (66.4) gives the velocity of propagation of the waves
from the known dependence of the frequency on the components of the
wave vector. This is a very important formula, which holds not only for
sound waves, but for all waves (for example, we have already applied it to
gravity waves in §12). We shall give here another derivation of this formula,
which puts in evidence the meaning of the velocity which it defines. Let us
consider a wave packet, which occupies some finite region of space. We
assume that its spectral composition includes monochromatic components
whose frequencies lie in only a small range; the same is true of the compo
nents of their wave vectors. Let to be some mean frequency of the wave
packet, and k a mean wave vector. Then, at some initial instant, the wave
packet is described by a function of the form
cf> = exp(zk.r)/(r). (66.8)
The function /(r) is appreciably different from zero only in a region which is
small (though it is large compared with the wavelength \jk). Its expansion as
a Fourier integral contains, by the above assumptions, components of the
form exp(av Ak), where Ak is small.
Thus each monochromatic component is, at the initial instant,
<£k = constantx exp[Y(k + Ak) • r] . (66.9)
The corresponding frequency is <o(k + Ak) (we recall that the frequency is a
function of the wave vector). Hence the same component at time t has the
form
<f> k = constantx exp[Y(k + Ak) • r — ico(k + Ak)*].
We use the fact that Ak is small, and expand w(k + Ak) in series, taking
only the first twO terms: co(k + Ak) = <x) + (dco/8k)'^k, where co = co(k) is
the frequency corresponding to the mean wave vector. Then <f> k becomes
<f> k = constant x exp>'(k r  cot)] expYAk' (r  tdco/dk)]. (66. 10)
If we now sum all the monochromatic components, with all the Ak that
occur in the wave packet, we see from (66.9) and (66.10) that the result is
<f> = exp[*(k • r  cot)]f(r  tdco/dk), (66.11)
where/is the same function as in (66.8). A comparison with (66.8) shows that,
after a time t, the amplitude distribution has moved as a whole through a
distance tdcojdk; the exponential coefficient of /in (66.11) affects only the
phase. Consequently, the velocity of the wave packet is
U = dcojdk. (66.12)
§67 Propagation of sound in a moving medium 259
This formula gives the velocity of propagation for any dependence of co
on k.f When a) = ck, with c constant, it of course gives the usual result
U = ojjk = c. In general, when co(k) is an arbitrary function, the velocity
of propagation is a function of the frequency, and the direction of propaga
tion may not be the same as that of the wave vector.
PROBLEM
Determine the altitude variation in the amplitude of sound propagated in an isothermal
atmosphere under gravity.
Solution. In an isothermal atmosphere (regarded as a perfect gas) the velocity of sound
is constant. The energy flux density evidently decreases along a ray in inverse proportion
to the square of the distance r from the source: cpv* ~ 1/r 2 . Hence it follows that the ampli
tude of the velocity fluctuations in the sound wave varies along a ray inversely as r\/ P ', according
to the barometric formula, p ~ exp(  figz/RT), where z is the altitude, ft the molecular weight
of the gas and R the gas constant.
§67. Propagation of sound in a moving medium
The relation w = ck between the frequency and the wave number is valid
only for a monochromatic sound wave propagated in a medium at rest. It is
not difficult to obtain a similar relation for a wave propagated in a moving
medium (and observed in a fixed system of coordinates).
Let us consider a homogeneous flow of velocity u. We take a fixed system
K of coordinates x, y, z, and also a system K' of coordinates x\ y\ z'
moving with velocity u relative to K. In the system K! the fluid is at rest,
and a monochromatic wave has the usual form cf> = constant x exp[/(k«r'  kct)].
The radius vector r' in the system K' is related to the radius vector r in
the system K by r' = rut. Hence, in the fixed system of coordinates, the
wave has the form <j> = constant x exp{*[k«r  (kc + k»u)*]}. The coeffici
ent of t in the exponent is the frequency w of the wave. Thus the frequency
in a moving medium is related to the wave vector k by
a, = ck + wk. (67.1)
The velocity of propagation is
dco/dk = ck/k + u; (67.2)
this is the vector sum of the velocity c in the direction of k and the velocity
u with which the sound is "carried along" by the moving fluid.
Using formula (67.1), we can investigate what is called the Doppler effect:
t The velocity defined by (66.12) is called the group velocity of the wave, and the ratio ufk the
phase velocity. However, it must be borne in mind that the phase velocity does not correspond to
any actual physical propagation.
Regarding the derivation given here it should be emphasised that the motion of the wave packet
without change of form (i.e. without change in the spatial distribution of the amplitude), expressed
by (66.11), is approximate, and results from the assumption that the range Ak is small. In general,
when U depends on <a, a wave packet is "smoothed out" during its propagation, and the region of
space which it occupies increases in size. It can be shown that the amount of this smoothing out
is proportional to the squared magnitude of the range Ak of the wave vectors which occur in the
composition of the wave packet.
260 Sound §67
the frequency of sound, as received by an observer moving relative to the
source, is not the same as the frequency of oscillation of the source.
Let sound emitted by a source at rest (relative to the medium) be received
by an observer moving with velocity u. In a system K' at rest relative to the
medium we have k = o> /c, where o> is the frequency of oscillation of the
source. In a system K moving with the observer, the medium moves with
velocity u, and the frequency of the sound is, by (67.1), a) = cku>k.
Introducing the angle 6 between the direction of the velocity u and that of
the wave vector k, and putting k = wqJc, we find that the frequency of the
sound received by the moving observer is
at = co [l(u/c) cos 6]. (67.3)
The opposite case, to a certain extent, is the propagation in a medium at
rest of a sound wave emitted from a moving source. Let u be now the velocity
of the source. We change from the fixed system of coordinates to a system K'
moving with the source; in the system K', the fluid moves with velocity u.
In K', where the source is at rest, the frequency of the emitted sound wave
must equal the frequency a> of the oscillations of the source. Changing the
sign of u in (67.1) and introducing the angle 6 between the directions of u
and k, we have o> = ck[l(u/c) cos ff]. In the original fixed system K,
however, the frequency and the wave vector are related by to = ck. Thus we
find
w = a>o/[l(ujc) cos 6]. (67.4)
This formula gives the relation between the frequency co of the oscillations
of a moving source and the frequency w of the sound heard by an observer
at rest.
If the source is moving away from the observer, the angle 6 between its
velocity and the direction to the observer lies in the range \n < 6 ^ it,
so that cos 6 < 0. It then follows from (67.4) that, if the source is moving
away from the observer, the frequency of the sound heard is less than co .
If, on the other hand, the source is approaching the observer, then
^ 8 < \n, so that cos 6 > 0, and the frequency to > o> increases with
u. For u cos 6 > c, according to formula (67.4) co becomes negative, which
means that the sound heard by the observer actually reaches him in the
reverse order, i.e. sound emitted by the source at any given instant arrives
earlier than sound emitted at previous instants.
As has been mentioned at the beginning of §66, the approximation of
geometrical acoustics corresponds to the case of small wavelengths, i.e. large
magnitudes of the wave vector. For this to be so the frequency of the sound
must in general be large. In the acoustics of moving media, however, the
latter condition need not be fulfilled if the velocity of the medium exceeds
that of sound. For in this case k can be large even when the frequency is
zero; from (67.1) we have for co = the equation
ck= uk, (67.5)
§67 Propagation of sound in a moving medium 261
and this has solutions if u > c. Thus, in a medium moving with supersonic
velocities, there can be steady small perturbations described (if k is sufficiently
large) by geometrical acoustics. This means that such perturbations are
propagated along rays.
Let us consider, for example, a homogeneous supersonic stream moving
with constant velocity u, whose direction we take as the #axis. The vector
k is taken to lie in the ryplane, and its components are related by
( M 2_ C 2)^2 = c 2 ky ^ (67.6)
which is obtained by squaring both sides of equation (67.5). To determine the
form of the rays, we use the equations of geometrical acoustics (66.4),
according to which x = 8wldk x ,y = dcojdky. Dividing one of these equations
by the other, we have dyjdx = (8a)l8ky)l(8o)l8k x ). This relation, however,
is, by the rule of differentiation for implicit functions, just the derivative
— dk x jdky taken at a constant frequency (in this case zero). Thus the equation
which gives the form of the rays from the known relation between k x and k y is
dyjdx = —dk x jdky. (67.7)
Substituting (67.6), we obtain
dyfdx = ±cJ\/{u 2 — c 2 ).
For constant u this equation represents two straight lines intersecting the
#axis at angles ± a, where sin a = cju.
We shall return to a detailed study of these rays in gas dynamics, where they
are very important; see in particular §§79, 96 and 109.
PROBLEMS
Problem 1. Derive an equation giving the form of sound rays propagated in a steadily
moving homogeneous medium with a velocity distribution u(x, y, z), when u <^ c every
where, f
Solution. Substituting (67.1) in (66.4), we obtain the equations of propagation of the
rays in the form
k = — (k'grad)u — kx curlu, r = v = ck/k + u.
Using these equations, and also
dufdt = du/dt + (v'grad)u = (vgrad)u a (c/£)(k«grad)u,
we calculate the derivative d(kv)fdt, retaining only terms as far as the first order in u. The
result is d(kv)ldt = —kvnX curlu, where n is a unit vector in the direction of v. But
d(kv)/dt = nd(kv)/dt +kv dnfdt. Since n and dn/dt are perpendicular (because n 2 = 1,
and therefore n«n = 0), it follows from the above equations that n = — nXcurl u. Intro
ducing the element of length along the ray dl = c dt, we can write finally
dn/dl =  n X curl u/c. (1)
This equation determines the form of the rays ; n is a unit tangential vector (and is no longer
in the same direction as k).
t It is assumed that the velocity u varies only over distances large compared with the wavelength
of the sound.
262 Sound §68
Problem 2. Determine the form of sound rays in a moving medium with a velocity distri
bution u x = u(z), u y = u t = 0.
Solution. Expanding equation (1), Problem 1, we find dn x /dl = (n z [c)du/dz, dn y /dl = 0;
the equation for n z need not be written down, since n a = 1. The second equation gives
n y = constant e= % )0 . In the first equation we write n z = dzjdl, and then we have by inte
gration n x = n Xt0 +u(z)lc These formulae give the required solution.
Let us assume that the velocity u is zero for z = and increases upwards (du/dz > 0).
If the sound is propagated "against the wind" (n x < 0), its path is curved upwards; if it is
propagated "with the wind" (n x > 0), its path is curved downwards. In the latter case a
ray leaving the point z = at a small angle to the xaxis (i.e. with n m , close to unity) rises
only to a finite altitude z = 2 max , which can be calculated as follows. At the altitude z mAX
the ray is horizontal, i.e. n z = 0. Hence we have
»z 2 + % 2 « n Xt o 2 + nyfi 2 + 2n x oulc = 1,
so that 2n x ,ou(z m& z)lc = w z ,o 2 , whence we can determine 2 max from the given function u(z)
and the initial direction n of the ray.
Problem 3. Obtain the expression of Fermat's principle for sound rays in a steadily moving
medium.
Solution. Fermat's principle is that the integral
cfkdl,
taken along a ray between two given points, is a minimum; k is supposed expressed as a
function of the frequency to and the direction n of the ray.f This function can be found by
eliminating v and k from the relations <o = ck+wh. and vn = ck/k+u. Fermat's principle
then takes the form
S^{V[(t 2 «2)d/2+(u.dl)2]U.dl}/(c2 tt 2) = 0.
In a medium at rest, this integral reduces to the usual one, j dljc.
§68. Characteristic vibrations
Hitherto we have discussed only oscillatory motion in infinite media, and
we have seen, in particular, that in such media waves of any frequency can be
propagated.
The situation is very different when we consider a fluid in a vessel of finite
dimensions. The equations of motion themselves (the wave equations) are
of course unchanged, but they must now be supplemented by boundary
conditions to be satisfied at the solid walls or at the free surface of the fluid.
We shall consider here only what are called free vibrations, i.e. those which
occur in the absence of variable external forces. Vibrations occurring as a
result of external forces are called forced vibrations.
The equations of motion for a finite fluid do not have solutions satisfying
the appropriate boundary conditions for every frequency. Such solutions
exist only for a series of definite frequencies <o. In other words, in a medium
of finite volume, free vibrations can occur only with certain frequencies.
These are called the characteristic frequencies of the fluid in the vessel
concerned.
The actual values of the characteristic frequencies depend on the size and
f See The Classical Theory of Fields, §71, Addison Wesley Press, Cambridge (Mass.) 1951.
§68 Characteristic vibrations 263
shape of the vessel. In any given case there is an infinite number of charac
teristic frequencies. To find them, it is necessary to examine the equations
of motion with the appropriate boundary conditions.
The order of magnitude of the first (i.e. smallest) characteristic frequency
can be seen at once from dimensional considerations. The only parameter
having the dimensions of length which appears in the problem is the linear
dimension / of the body. Hence it is clear that the wavelength Ai correspond
ing to the first characteristic frequency must be of the order of /, and the order
of magnitude of the frequency cui itself is obtained by dividing the velocity
of sound by the wavelength. Thus
Ai  /, oil ~ c\l. (68.1)
Let us ascertain the nature of the motion in characteristic vibrations.
If we seek a solution of the wave equation for the velocity potential (say)
which is periodic in time, of the form <j> — <j>q{x, y, z)e~ i<ot , then we have for
<f>o the equation
A<h + («> 2 lc 2 )<h = 0. (68.2)
In an infinite medium, where no boundary conditions need be applied, this
equation has both real and complex solutions. In particular, it has a solution
proportional to e ik ' T , which gives a velocity potential of the form
<j> = constant xexp[i(k'r—cot)].
Such a solution represents a wave propagated with a definite velocity — a
travelling wave.
For a medium of finite volume, however, complex solutions cannot in
general exist. This can be seen as follows. The equation satisfied by <f>Q
is real, and the boundary conditions are real also. Hence, if <f>o(x, y, z) is a
solution of the equations of motion, the complex conjugate function <f>o*
is also a solution. Since, however, the solution of the equations for given
boundary conditions is in general uniquef apart from a constant factor, we
must have <£o* = constant x <£o, where the constant is complex and its
modulus is clearly unity. Thus <f>o must be of the form <f>o = f(x, y, z)e~ ia ,
the function / and the constant a being real. The potential <f> is thus of the
form (taking the real part of (f>oe~ iwt )
<f> = f{ x >y> % ) cos(cot + <x), (68.3)
i.e. it is the product of some function of the coordinates and a simple periodic
function of the time.
This solution has properties entirely different from those of a travelling
wave. In the latter, where <f> = constant xcos(k«r— «j^ + a), the phase
k«r — cot +<x of the oscillations at different points in space is different at any
given instant, except only at points for which k • r differs by an integral
f This may not be true when the vessel is highly symmetrical in form (e.g. a sphere).
264 Sound §68
multiple of the wavelength. In the wave represented by (68.3), all points are
oscillating in the same phase tot + at. at any given instant. Such a wave is
obviously not "propagated" ; it is called a stationary wave. Thus the charac
teristic vibrations are stationary waves.
Let us consider a stationary plane sound wave, in which all quantities are
functions of one coordinate only (x, say) and of time. Writing the general
solution of 8 2 <f>ojdx 2 + 2 co(f)olc 2 = in the form <f>o = a cos(a»*/c+/S), we have
<f> = a cos(cor+a) cos(cox/c+l3). By an appropriate choice of the origins
of x and t, we can make a and /S zero, so that
<f> = a cos cot cos cox/c. (68.4)
For the velocity and pressure in the wave we have
v = 8(f>Jdx = — (aco/c) cos cot sin cox/c ;
p' = —p 8<f>/8t = pco sin cot cos cox/c.
At the points x = 0, ire/ to, 2,nc/to, ..., which are at a distance ttc/co = £A
apart, the velocity v is always zero ; these points are called nodes of the velocity.
The points midway between them (x = ttc/Ico, Zttc/2co, ...) are those at
which the amplitude of the time variations of the velocity is greatest. These
are called antinodes. The pressure p' evidently has nodes and antinodes in
the reverse positions. Thus, in a stationary plane wave, the nodes of the
pressure are the antinodes of the velocity, and vice versa.
An interesting case of characteristic vibrations is that of the vibrations of
a gas in a vessel having a small aperture (a resonator). In a closed vessel the
smallest characteristic frequency is, as we know, of the order of c/l, where /
is the linear dimension of the vessel. When there is a small aperture, however,
new characteristic vibrations of considerably smaller frequency appear.
These are due to the fact that, if there is a pressure difference between the
gas in the vessel and that outside, this difference can be equalised by the
motion of gas into or out of the vessel. Thus oscillations appear which
involve an exchange of gas between the resonator and the outside medium.
Since the aperture is small, this exchange takes place only slowly, and hence
the period of the oscillations is large, and the frequency correspondingly
small (see Problem 2). The frequencies of the ordinary vibrations occurring
in a closed vessel are practically unchanged by the presence of a small aper
ture.
PROBLEMS
Problem 1 . Determine the characteristic frequencies of sound waves in a fluid contained
in a cuboidal vessel.
Solution. We seek a solution of the equation (68.2) in the form
</>q = constant x cos qx cos ry cos sz,
where q 2 +r i +s* = w 2 /c 2 . At the walls of the vessel we have the conditions v x = d<f>/dx =
for x = and a, d<f>[dy = for y = and b, 8<f>/dz = for z = and c, where a, b, c are
the sides of the cuboid. Hence we find q — mnla, r = nn/b, s = pn\c, where m, n, p are
any integers. Thus the characteristic frequencies are
co 2 = c 2 7T 2 (m 2 /a 2 +n 2 /b 2 +p 2 /c 2 ).
§69 Spherical waves 265
Problem 2. A narrow tube of crosssectional area S and length / is fixed to the aperture of
a resonator. Determine the characteristic frequency.
Solution. Since the tube is narrow, in considering oscillations accompanied by the
movement of gas into and out of the resonator we can suppose that only the gas in the tube
has an appreciable velocity, while the gas in the vessel is almost at rest. The mass of gas in
the tube is Spl, and the force on it is S(p —p), where p and p are the gas pressures inside and
outside the resonator respectively. Hence we must have Spiv = S(p—p ), where v is the
gas velocity in the tube. The time derivative of the pressure is given by p = c 2 p, and the
decrease per unit time in the gas density in the resonator ( — p) can be supposed equal to the
mass of gas leaving the resonator per unit time (Spv) divided by the volume V of the resonator.
Thus we have£ «= —c^SpvfV, whence
/>"= c 2 SpvjV = c*S(pp )llV.
This equation gives p—po = constant X cos to t, where the characteristic frequency
(o — c\Z{SjlV). This is small compared with cjL (where L is the linear dimension of the
vessel), and the wavelength is therefore large compared with L.
In solving this problem we have supposed that the linear amplitude of the oscillations of
gas in the tube is small compared with its length /. If this were not so, the oscillations would
be accompanied by the outflow of a considerable fraction of the gas in the tube, and the linear
equation of motion used above would be inapplicable.
§69. Spherical waves
Let us consider a sound wave in which the distribution of density, velocity,
etc., depends only on the distance from some point, i.e. is spherically sym
metrical. Such a wave is called a spherical wave.
Let us determine the general solution of the wave equation which represents
a spherical wave. We take the wave equation for the velocity potential:
/\<f> — (llc 2 )d 2 <f>jdt 2 = 0. Since ^ is a function only of the distance r from the
centre and of the time t y we have, using the expression for the Laplacian in
spherical coordinates,
8 2 (f> 1 d l 86 \
Z = C 2 ( r 2_Z . (69.1)
dt 2 r 2 dr \ dr J
We seek a solution in the form <f> = /(r, t)jr. Substituting, we have after
a simple calculation the following equation for /: d 2 fjdt 2 = c 2 d 2 fjdr 2 . This is
just the ordinary onedimensional wave equation, with the radius r as
the coordinate. The solution of this equation is, as we know, of the form
f = fi(ct—r)+f2(ct+r), where /i and f 2 are arbitrary functions. Thus the
general solution of equation (69.1) is of the form
^.^ =! ) + ^+o (69 . 2)
r r
The first term is an outgoing wave, propagated in all directions from the origin.
The second term is a wave coming in to the centre. Unlike a plane wave,
whose amplitude remains constant, a spherical wave has an amplitude which
decreases inversely as the distance from the centre. The intensity in the wave
is given by the square of the amplitude, and falls off inversely as the square of
the distance, as it should, since the total energy flux in the wave is distributed
over a surface whose area increases as r 2 .
266 Sound §69
The variable parts of the pressure and density are related to the potential
by p' = —pd(j>jdt, p = —{pjc 2 )d<j>Jdt, and their distribution is determined by
formulae of the same form as (69.2). The (radial) velocity distribution, how
ever, being given by the gradient of the potential, is of the form
v = y f) ; /a(rf+f) ). (69.3)
If there is no source of sound at the origin, the potential (69.2) must remain
finite for r = 0. For this to be so we must have/i(c£) = —fact), i.e. <f> is
of the form
f(ctr)f(ct+r)
<f> = (69.4)
r
(a stationary spherical wave). If there is a source at the origin, on the other
hand, the potential of the outgoing wave from it is <f> = f(ct—r)/r; it need
not remain finite at r = 0, since the solution holds only for the region
outside sources.
A monochromatic stationary spherical wave is of the form
sinkr
<£ = Ae~M , (69.5)
r
where k = wjc. An outgoing monochromatic spherical wave is given by
<f> = Ae^^lr. (69.6)
It is useful to note that this expression satisfies the differential equation
A<f> + &<j> = ^rAe^B(r), (69.7)
where on the righthand side we have the delta function S(r) = 8(x)B(y)8(z).
For S(r) = everywhere except at the origin, and we return to the homo
geneous equation (69.1); and, integrating (69.7) over the volume of a small
sphere including the origin (where the expression (69.6) reduces to Ae^^/r)
we obtain — 4irAe~ i<ot on each side.
Let us consider an outgoing spherical wave, occupying a spherical shell
outside which the medium is either at rest or very nearly so ; such a wave can
originate from a source which emits during a finite interval of time only, or
from some region where there is a sound disturbance (cf. the end of §71,
and §73, Problem 4). Before the wave arrives at any given point, the potential
is <j> = 0. After the wave has passed, the motion must die away; this means
that <f> must become constant. In an outgoing spherical wave, however, the
potential is a function of the form <j> = f(ct — r)/r; such a function can tend
to a constant only if the function / is zero identically. Thus the potential
must be zero both before and after the passage of the wave.f From this we
t Unlike what happens for a plane wave, after whose passage we can have ^ = constant =£
§69 Spherical waves 267
can draw an important conclusion concerning the distribution of conden
sations and rarefactions in a spherical wave.
The variation of pressure in the wave is related to the potential by
p' = — pd<f}{dt. From what has been said above, it is clear that, if we integrate
p' over all time for a given r, the result is zero :
00
JVd* = 0. (69.8)
—00
This means that, as the spherical wave passes through a given point, both
condensations (p' > 0) and rarefactions (/>' < 0) will be observed at that
point. In this respect a spherical wave is markedly different from a plane
wave, which may consist of condensations or rarefactions only.
A similar pattern will be observed if we consider the manner of variation of
p' with distance at a given instant; instead of the integral (69.8) we now
consider another which also vanishes, namely
00
jrp'dr = 0. (69.9)
o
PROBLEMS
Problem 1. At the initial instant, the gas inside a sphere of radius a is compressed so that
p' = constant = A; outside this sphere, p' = 0. The initial velocity is zero in all space.
Determine the subsequent motion.
Solution. The initial conditions on the potential are <f> — for t = 0, and r < a or
r > a; <j> = F(r) for t = 0, where F(r) = for r > a and F(r) = c 2 A/p for r < a. We
seek $ in the form
xi * f{ctr)f{ct + r)
r
From the initial conditions we obtain /( — r) — /(r) = 0,f'(—r)—f'(r) = rF(r)jc. From the
first equation we have f'(—r)+f'(r) = 0, which together with the second equation gives
/'(r) = —/'(—*)— —rF(r)/2c. Finally, substituting the value of F(r), we find the following
expressions for the derivative /'(f) and the function /(f) itself:
for  > a, /'(*) = 0, fit) = 0;
for III < a, /'() = cZbfo, /() = <2_ a 2 )A/4/>j
which give the solution of the problem. If we consider a point with r > a, i.e. outside the
region of the initial compression, we have for the density
for t < (r — a)jc, p = 0;
for (ra)/c < t < (r + a)jc, p = %(rct)&{r;
for t > (r+a)lc, p = 0.
The wave passes the point considered during a time interval 2a/c; in other words, the wave
has the form of a spherical shell of thickness 2a, which at time t lies between the spheres of
radii ct—a and ct+a. Within this shell the density varies linearly; in the outer part (r > ct),
the gas is compressed (p' > 0), while in the inner part (r < ct) it is rarefied (p' < 0).
268 Sound §70
Problem 2. Determine the characteristic frequencies of centrally symmetrical sound
oscillations in a spherical vessel of radius a.
Solution. From the boundary condition d<j>jdr = f or r = a (where <f> is given by (69.5))
we find tan ka = ha, which determines the characteristic frequencies. The first (lowest)
frequency is a>x = 4*49 c\a.
§70. Cylindrical waves
Let us now consider a wave in which the distribution of all quantities is
homogeneous in some direction (which we take as the araxis) and has com
plete axial symmetry about that direction. This is called a cylindrical wave,
and in it we have <f> = <f>(R, t), where R denotes the distance from the #axis.
Let us determine the general form of such an axisymmetric solution of the
wave equation. This can be done by starting from the general spherically
symmetrical solution (69.2). R is related to r by r 2 = R 2 + z 2 , so that <f> as
given by formula (69.2) depends on z when R and t are given. A function
which depends on R and t only and still satisfies the wave equation can be
obtained by integrating (69.2) over all z from — oo to oo, or equally well
from to oo. We can convert the integration over z to one over r. Since
z = \/(r 2 i? 2 ), dz = r drj<\/(r 2 — R. 2 ). When z varies from to oo, r
varies from R to oo. Hence we find the general axisymmetric solution to be
r fi(ctr) r fa(ct + r) t
j, = jL± f_dr+ — —dr y (70.1)
V(r*R 2 ) J ^/(r 2 R 2 )
where f\ and fa are arbitrary functions. The first term is an outgoing cylin
drical wave, and the second an ingoing one.
Substituting in these integrals ct±r — g, we can rewrite formula (70.1) as
ctR
(• m* + f mw , (702)
oo ct+R
We see that the value of the potential at time t in the outgoing cylindrical
wave is determined by the values of fa at times from — oo to t — Rjc;
similarly, the values of fa which affect the ingoing wave are those at times
from t + RIc to infinity.
As in the spherical case, stationary waves are obtained when/i(£) = —fa(£).
It can be shown that a stationary cylindrical wave can also be represented in
the form
ct+R
9 J V[R 2 (tctf] ( '
ctR
where F(g) is another arbitrary function.
§70 Cylindrical waves 269
Let us derive an expression for the potential in a monochromatic cylindrical
wave. The wave equation for the potential <f)(R, t) in cylindrical coordinates
is
R 8R\ 8R/ c*
8U
— = 0.
dp
In a monochromatic wave <f> = e* w< /(i?), and we have for the function /(i?)
the equation/" +/'/!?+ A 2 / = 0. This is Bessel's equation of order zero.
In a stationary cylindrical wave, <f> must remain finite for R = 0; the appro
priate solution isJo(kR), where Jo is a Bessel function of the first kind. Thus,
in a stationary cylindrical wave,
<f> = Aei»J (kR). (70.4)
For R — the function Jo tends to unity, so that the amplitude tends to the
finite limit A. At large distances R, Jo may be replaced by its asymptotic
expression, and <£ then takes the form
<k = A / iV***. (70.5)
The solution corresponding to a monochromatic outgoing travelling wave is
<£ = Ae***H<P(kR) y (70.6)
where Ho {1) is the Hankel function of the first kind, of order zero. For
R > this function has a logarithmic singularity:
<f> £ (2^/tt) log(£i?)<?K (70.7)
At large distances we have the asymptotic formula
i2exp[i(kRa>t~frr)]
V(kR)
We see that the amplitude of a cylindrical wave diminishes (at large distances)
inversely as the square root of the distance from the axis, and the intensity
therefore decreases as ljR. This result is obvious, since the total energy
flux is distributed over a cylindrical surface, whose area increases propor
tionally to R as the wave is propagated.
A cylindrical outgoing wave differs from a spherical or plane wave in the
important respect that it has a forward front but no backward front: once the
sound disturbance has reached a given point, it does not cease, but diminishes
comparatively slowly as t > oo. Suppose that the function /i(£) in the
first term of (70.2) is different from zero only in some finite range
h < £ < h Then, at times such that ct > R+ &> we have
S2
Mm
270 Sound §71
As t > oo, this expression tends to zero as
= Vt j* 1 ® 6 **
Si
i.e. inversely as the time.
Thus the potential in an outgoing cylindrical wave, due to a source which
operates only for a finite time, vanishes, though slowly, as t > oo. This
means that, as in the spherical case, the integral of p' over all time is zero:
00
fp'dt = 0. (70.9)
—oo
Hence a cylindrical wave, like a spherical wave, must necessarily include both
condensations and rarefactions.
§71. The general solution of the wave equation
We shall now derive a general formula giving the solution of the wave
equation in an infinite fluid for any initial conditions, i.e. giving the velocity
and pressure distribution in the fluid at any instant in terms of their initial
distribution.
We first obtain some auxiliary formulae. Let <f>(x, y, z, t) and i[j(x, y, z, t)
be any two solutions of the wave equation which vanish at infinity. We
consider the integral
/=/ («^#)dF,
taken over all space, and calculate its time derivative. Since <f> and ifj satisfy
the equations A<f>$lc 2 = and A«A~$/c 2 = 0, we have
dl/dt = j {<f4^)dV = c* j (M«AM<£)dF
= c 2 J div(<£ grad ifj — if/ grad <f>)dV.
The last integral can be transformed into an integral over an infinitely distant
surface, and is therefore zero. Thus we conclude that dljdt = 0, i.e. J is
independent of time :
f {<j4  #)d V = constant. (71.1)
Next, let us consider the following particular solution of the wave
equation :
4> = 8[rc{tot)]lr (71.2)
(where r is the distance from some given point O, to is some definite instant,
§71 The general solution of the wave equation 271
and 8 denotes the delta function), and calculate the integral of ifj over all
space. We have
00 00
j l fsdV = J* if*4irr 2 dr = 4tt j rS[rc(t t)]dr.
o o
The argument of the delta function is zero for r = c(to — t) (we assume that
*o > *)• Hence, from the properties of the delta function, we find
jt/,dV = +nc(tot). (71.3)
Differentiating this equation with respect to time, we obtain
ffdV = 4ttc. (71.4)
We now substitute for «/r, in the integral (71.1), the function (71.2), and
take <f> to be the required general solution of the wave equation. According
to (71.1), J is a constant; using this, we write down the expressions for /
at the instants t = and t = to, and equate the two. For t = t o the two
functions t[t and if, are each different from zero only for r = 0. Hence, on
integrating, we can put r = in (f> and (f> (i.e. take their values at the point O),
and take <f> and <f> outside the integral:
I = <f>{x,y, z y t ) j 4 &V <f>{x,y, z, t ) j tfi dV,
where x t y, z are the coordinates of O. According to (71.3) and (71.4),
the second term is zero for t — to, and the first term gives
I =  47TC(j}(x > y, z y to).
Let us now calculate /for t = 0. Putting = dif/Jdt = —dtpjdtoy and
denoting by <£o the value of the function <f> for t = 0, we have
/= _ (L^ + ^AdV^ £ (foifttodV f^MidF.
We write the element of volume as dV = r 2 drdo, where do is an element of
solid angle, and then we obtain, by the properties of the delta function,
J foifftodV = j cf>orh(rcto)drdo = ct j <f>Q r = cU do\
the integral of <£o«A is similar. Thus
a
dt
(cto <£o,rct<,do) — Ct <f>0,r=ct do.
272 Sound §71
Finally, equating the two expressions for / and omitting the suffix zero in to,
we obtain
<l>(x,y, *> = 7 7C h,r=ct do) + 1\ <f> 0t r==ct do . (71.5)
This formula, called Poisson's formula, gives the spatial distribution of
the potential at any instant in terms of the distribution of the potential and
its time derivative (or, equivalently, in terms of the velocity and pressure
distribution) at some initial instant. We see that the value of the potential
at time t is determined by the values of <j> and (j> at time t = on the surface
of a sphere centred at O, of radius ct.
Let us suppose that, at the initial instant, <f>o and <j>o are different from zero
only in some finite region of space, bounded by a closed surface C (Fig. 34).
Fig. 34
We consider the values of <f> at subsequent instants at some point O. These
values are determined by the values of <£o and <£o at a distance ct from O.
The spheres of radius ct pass through the region within the surface C only
for djc < t ^ Djc, where d and D are the least and greatest distances from
the point O to the surface C. At other instants, the integrands in (71.5)
are zero. Thus the motion at O begins at time t = djc and ceases at time
t = Djc. The wave propagated from the region inside C has a forward
front and a backward front. The motion begins when the forward front
arrives at the point in question, while on the backward front particles pre
viously oscillating come to rest.
PROBLEM
Derive the formula giving the potential in terms of the initial conditions for a wave depend
ing on only two coordinates, * and y.
Solution. An element of area of a sphere of radius ct can be written d/ = c 2 t 2 do, where do
is an element of solid angle. The projection of d/ on the xyplane i s d* dy = df\/[(ct) 2 — p s ]fct,
where p is the distance of the point x, y from the centre of the sphere. Comparing the two
expressions, we can write do = da: dy/c£\/[(c£) 2 ~P 2 ] Denoting by x, y the coordinates of
the point where we seek the value of ^, and by i, 17 the coordinates of a variable point in the
region of integration, we can therefore replace do in the general formula (71.5) by
d£ dt]/ct\/[(ct) 2 — (x— i) 2 — (y— i?) 2 ], doubling the resulting expression because d* dy is the
§72
The lateral wave
273
projection of two elements of area on opposite sides of the #yplane. Thus
V[(ct) 2 (xo 2 (yv) 2 ]
IttC J J
+
^o(£*?)d£d7?
2ncJJ V[(ct) 2 (x0 2 (yn) 2 ]'
where the integration is over a circle centred at O, of radius ct. If <j> and <f> are zero except
in a finite region C of the xyplane (or, more exactly, except in a cylindrical region with its
generators parallel to the saxis), the oscillations at the point O (Fig. 34) begin at time
t = djc, where d is the least distance from O to a point in the region. After this time, however,
circles of radius ct > d centred at O will always enclose part or all of the region C, and <f>
will tend only asymptotically to zero. Thus, unlike threedimensional waves, the two
dimensional waves here considered have a forward front but no backward front (cf. §70).
Reflected wave
Fig. 35
§72. The lateral wave
The reflection of a spherical wave from the surface separating two media
is of particular interest in that it may be accompanied by an unusual pheno
menon, the appearance of a lateral wave.
Let Q (Fig. 35) be the source of a spherical sound wave in medium 1, at a
274 Sound §72
distance / from the infinite plane surface separating media 1 and 2. The
distance / is arbitrary, and need not be large compared with the wavelength A.
Let the densities of the two media be pi, pz, and the velocities of sound in
them ci, cz. We suppose first that c\ > ci ; then, at distances from the source
large compared with A, the motion in medium 1 will be a superposition of
two outgoing waves. One of these is the spherical wave emitted by the source
(the direct wave) ; its potential is
<£i° = eMr/r, (72.1)
where r is the distance from the source, and the amplitude is arbitrarily taken
to be unity. We shall, for brevity, omit the factor e~ t<ot from all expressions
in the present section.
Fig. 36
The wave surfaces of the second (reflected) wave are spheres centred at Q\
the image of the source Q in the plane of separation; this is the locus of
points P reached at a given time by rays which leave Q simultaneously and
are reflected from the plane in accordance with the laws of geometrical acou
stics (in Fig. 36, the ray QAP with angles of incidence and reflection 6 is
shown). The amplitude of the reflected wave decreases inversely as the
distance r' from the point Q' (which is sometimes called an imaginary
source), but depends also on the angle 6, as if each ray were reflected with the
coefficient corresponding to the reflection of a plane wave at the given angle
of incidence 6. In other words, at large distances the reflected wave is given
by the formula
, = *** r p 2g2  m V( gl 2 g2 2 sin 2g) 
r' Lp2t2 + piV(^i 2 ^ 2 sin26l)J' K ' j
cf. formula (65.4) for the reflection coefficient for a plane wave. This formula,
which is clearly valid for large r', can be rigorously derived by the method
shown below.
A more interesting case is that where c\ < c<l. Here, besides the ordinary
reflected wave (72.2), another wave appears in the first medium. The
§72 The lateral wave 275
chief properties of this wave can be seen from the following simple con
siderations.
The ordinary reflected ray QAP (Fig. 36) obeys Fermat's principle in the
sense that it is the quickest path from Q to P, among paths lying entirely in
medium 1 and involving a single reflection. When c\ < C2, however, Fermat's
principle is also satisfied by another path, where the ray is incident on the
boundary at the critical angle of total internal reflection #o (sin #o = ^1/^2),
then is propagated in medium 2 along the boundary, and finally returns to
medium 1 at the angle do. The path is QBCP in Fig. 36, and it is evident
that 9 > 6q. It is easy to see that this path also has the extremal property:
the time taken to traverse it is less than for any other path from Q to P lying
partly in medium 2.
The geometrical locus of points P reached at the same time by rays which
simultaneously leave Q along the path QB, and then return to medium 1 at
various points C, is evidently a conical surface whose generators are perpen
dicular to lines drawn from the imaginary source Q' at an angle 6q.
Thus, if c\ < C2, together with the ordinary reflected wave, which has
a spherical front, there is propagated in medium 1 another wave, which has a
conical front extending from the plane of separation (where it meets the
refracted wave front in medium 2) to the point where it touches the spherical
front of the reflected wave ; this occurs along the line of intersection with a
cone of semiangle #o and axis QQ' (Fig. 35). This conical wave is called the
lateral wave.
It is easy to see by a simple calculation that the time along the path QBCP
(Fig. 36) is less than along the path QAP to the same point P. This means
that a sound signal from the source Q reaches an observer at P first as the
lateral wave, and only later as the ordinary reflected wave.
It must be borne in mind that the lateral wave is an effect of wave acoustics,
despite the fact that it allows the above simple interpretation in terms of the
concepts of geometrical acoustics. We shall see below that the amplitude of
the lateral wave tends to zero in the limit A > 0.
Let us now make a quantitative calculation. The propagation of a mono
chromatic sound wave from a point source is described by equation (69.7) :
A<f> + k*<f> = 47rS(rl), (72.3)
where k = wjc and 1 is the radius vector of the source. The coefficient
of the delta function is chosen so that the direct wave has the form (72.1).
In what follows we take a system of coordinates with the ryplane as the
plane of separation and the #axis along QQ', with the first medium in z > 0.
At the surface of separation the pressure and the ^component of the velocity,
or (equivalently) p<f> and d(f>/dz, must be continuous.
Using the general Fourier method, we obtain the solution in the form
co 00
<f> = j j <f> K (z) txp[i( Kx x+ Ky y)] &K X &Ky, (72.4)
276 Sound §72
where
1 00 00
^*) = 4^2 J J ^expt'Va^+^Ky)]^^. ( 72  5 )
—oo — oo
From the symmetry relative to the xyplane it is evident that <£ K can depend
only on the quantity \k\ = \/(k x 2 + k v 2 ). Using the wellknown formula
1 r
Jo(u) = — cos(w sin <£)d<£,
2tt J
o
we can therefore write (72.4) as
<f> = 2tt j<f> K (z)J (KR) K dK, (72.6)
o
where R = \/(x 2 +y 2 ) is the cylindrical coordinate (the distance from the
jsraxis). It is convenient for the subsequent calculations to transform this
formula into one in which the integral is taken from — oo to oo, expressing
the integrand in terms of the Hankel function Hq {1 \u). The latter has a
logarithmic singularity at the point u = 0; if we agree to go from positive
to negative real u by passing above the point u = in the complex wplane,
then Ho (1) (u) = H ^\ue in ) = Ho (1 \u)2J (u). Using this relation, we
can rewrite (72.6) as
<f> = 7t J <f> K (z)Ho a \KR) K die. (72.7)
—00
From equation (72.3) we find for the function j> K the equation
d*2
/ co 2 \ 1
(k 2 )^= 8(*/). (72.8)
The delta function on the righthand side of the equation can be eliminated
by imposing on the function <f> K (z) (satisfying the homogeneous equation)
the boundary conditions at z = I:
[d^/d^[d^/d^_= 1/77.
The boundary conditions at z = are
l>^]o+[>^]o = 0,
[d<^d*]o + [d4/d*]o = 0.
§72 The lateral wave 277
We seek a solution in the form
<f> K = AeW for z > I, \
<f> K = Be^ + Ce^ for / > z > 0, [ (72.11)
<f> K = De^ z for > z. J
Here
^2 = K 2k^, [M2 2 = K 2 k 2 2 (ki = Qi\c\, k 2 = C0/C 2 ),
and we must put
u, = +VU 2 A 2 )for k > k,
^ VV } (72.12)
/x = — i\/(k 2 — k 2 ) for k < k.
The first of these is necessary so that <j> should not increase without limit as
z > oo, and the second so that <f> should represent an outgoing wave. The
conditions (72.9) and (72.10) give four equations which determine the co
efficients A, B, C and D. A simple calculation gives
B = c ^wpm c = e l/l1
D=C ^1_, A = B + Ce 2 ^.
/*lp2 + )U,2pi
(72.13)
For p2 = pi, c 2 = c\ (i.e. when all space is occupied by one medium),
B is zero and A = Ce 21 ^ ; the corresponding term in <f> is evidently the dire ct
wave (72.1), and the reflected wave in which we are interested is therefore
oo
cf>i' = tt j B(K)e^W»(KR)KdK. (72.14)
—oo
In this expression the path of integration has to be specified. It passes
above the singular point k = (in the complex /cplane), as we have already
mentioned. The integrand also has singular points (branch points) at
k = ±ki, ± k 2 , where m or \i 2 vanishes. In accordance with the conditions
(72.10), the contour must pass below the points +ki, +k 2 , and above the
points — ki, —k 2 .
Let us investigate the resulting expression for large distances from the
source. Replacing the Hankel function by its wellknown asymptotic
expression, we obtain
<£i' = f f " lp2 ~^ 2pl /JL^pifr+^ + iKRidK. (72.15)
J M^1P2 + f*2pl) N LinR
Fig. 37 shows the path of integration C for the case c\ > c%. The integral can
be calculated by means of the saddlepoint method. The exponent
i[(z + l)\/(ki 2 — k 2 ) + kR] has an extremum at the point where
k/V(&i 2 k 2 ) = R/(z + l) = r' sind/r' cosd = tanfl,
278
Sound
§72
i.e. k = k\ sin #, where 6 is the angle of incidence (see Fig. 35). On changing
to the path of integration C" which passes through this point at an angle of
7r/4 to the axis of abscissae, we obtain formula (72.2).
Fig. 37
In the case ci < c^ (i.e. h\ > £2), the point k = k\ sin 6 lies between
&2 and ki if sin 6 > hi\k\ = c\\c^ — sin do, ie. if 6 > 6q (Fig. 38). In this
case the contour C" must make a loop round the point kz, and we have,
besides the ordinary reflected wave (72.2), a wave <f>i" given by the integral
(72.15) taken around the loop, which we call C" . This is the lateral wave.
The integral is easily calculated if the point &i sin 9 is not close to hi, i.e.
if the angle 6 is not close to the internalreflection angle 0o«t
Fig. 38
Near the point k = &2, ft2 is small ; we expand the coefficient of the expo
nential in the integrand of (72.15) in powers of fi2. The zeroorder term has
no singularity at k = k%, and its integral round C" is zero. Hence we have
#>= _ f^L /_^ X p[_(*+/) / , 1+ ;, fj R]dK.
J MrP2 V 2&77T
c
(72.16)
Expanding the exponent in powers of k — k% and integrating round the loop
t For an investigation of the lateral wave for all values of d, see L. Brekhovskikh, Zhurnal tekh
nicheskoi fiziki 18, 455, 1948. This paper gives also the next term in the expansion of the ordinary
reflected wave in powers of XjR. We may mention here that, for angles d close to (in the case
c 1 < c 2 ), the ratio of the correction term to the leading term falls off with distance as (A/i?)i, and
not as X/R.
§73 The emission of sound 279
C", we have after a simple calculation the following expression for the poten
tial of the lateral wave :
<k " = 2*Pifoexp[iftir'cos(e fl)]
^ r'a/o^iVtcos^osin^sinS^o^)]' '
In accordance with the previous results, the wave surfaces are the cones
r' cos(d—9o) = R sin 6o + (z+l) cos do = constant. In a given direction, the
wave amplitude decreases inversely as the square of the distance r'. We see
also that this wave disappears in the limit A > 0. For > do, the expres
sion (72.17) ceases to be valid; in actual fact, the amplitude of the lateral
wave in this range of 6 decreases with distance as r' 5/4 .
§73. The emission of sound
A body oscillating in a fluid causes a periodic compression and rarefaction
of the fluid near it, and thus produces sound waves. The energy carried
away by these waves is supplied from the kinetic energy of the body. Thus
we can speak of the emission of sound by oscillating bodies. f
In the general case of a body of arbitrary shape oscillating in any manner,
the problem of the emission of sound waves must be solved as follows. We
take the velocity potential <f> as the fundamental quantity; it satisfies the wave
equation
A^G/^W/ 3 ' 2 = 0 (73.1)
At the surface of the body, the normal component of the fluid velocity must
be equal to the corresponding component of the velocity u of the body:
8<f>/8n = u n . (73.2)
At large distances from the body, the wave must become an outgoing spherical
wave. The solution of equation (73.1) which satisfies these boundary con
ditions and the condition at infinity determines the sound wave emitted by
the body.
Let us consider the two boundary conditions in more detail. We suppose
first that the frequency of oscillation of the body is so large that the length
of the emitted wave is very small compared with the dimension / of the body:
A < I (73.3)
In this case we can divide the surface of the body into portions whose dimen
sions are so small that they may be approximately regarded as plane, but
yet are large compared with the wavelength. Then we may suppose that each
f In what follows we shall always suppose that the velocity u of the oscillating body is small com
pared with the velocity of sound. Since u ~ aco, where a is the linear amplitude of the oscillations,
this means that a <^ A.
The amplitude of the oscillations is in general supposed small in comparison with the dimensions
of the body also, since otherwise we do not have potential flow near the body (cf. §9). This con
dition is unnecessary only for pure pulsations, when the solution (73.7) used below is really a direct
deduction from the equation of continuity.
280 Sound §73
such portion emits a plane wave, in which the fluid velocity is simply the
normal component u n of the velocity of that portion of the surface. But the
mean energy flux in a plane wave is (see §64) cpv 2 , where v is the fluid velocity
in the wave. Putting v — u n and integrating over the whole surface of the
body, we reach the result that the mean energy emitted per unit time by
the body in the form of sound waves, i.e. the total intensity of the emitted
sound, is
I = c P j>u n zdf. (73.4)
It is independent of the frequency of the oscillations (for a given velocity
amplitude).
Let us now consider the opposite limiting case, where the length of the
emitted wave is large compared with the dimension of the body:
A > I (73.5)
Then we can neglect the term (llc 2 )d 2 <f>[dt 2 , in the general equation (73.1),
near the body (at distances small compared with the wavelength). For this
term is of the order of cd 2 ^/c 2 ~ (f>[X 2 , whereas the second derivatives with
respect to the coordinates are, in this region, of the order of </>// 2 .
Thus the flow near the body satisfies Laplace's equation, A<£ = 0. This
is the equation for potential flow of an incompressible fluid. Consequently
the fluid near the body moves as if it were incompressible. Sound waves
proper, i.e. compression and rarefaction waves, occur only at large distances
from the body.
At distances of the order of the dimension of the body and smaller, the
required solution of the equation /\<j> = cannot be written in a general form,
but depends on the actual shape of the oscillating body. At distances large
compared with /, however (though still small compared with A, so that the
equation /\<j> = remains valid), we can find a general form of the solution
by using the fact that <f> must decrease with increasing distance. We have
already discussed such solutions of Laplace's equation in §11. As there,
we write the general form of the solution as
cj> = (a/r)+A.grad(l/r), (73.6)
where r is the distance from an origin anywhere inside the body. Here, of
course, the distances involved must be large compared with the dimension
of the body, since we cannot otherwise restrict ourselves to the terms in <j>
which decrease least rapidly as r increases. We have included both terms in
(73.6), although it must be borne in mind that the first term is sometimes
absent (see below).
Let us ascertain in what cases this term — a/r is nonzero. We found
in §11 that a potential —a\r results in a nonzero value Airpa of the mass flux
through a surface surrounding the body. In an incompressible fluid, how
ever such a mass flux can occur only if the total volume of fluid enclosed within
§73 The emission of sound 281
the surface changes. In other words, there must be a change in the volume
of the body, as a result of which the fluid is either expelled from or "sucked
into" the volume of space concerned. Thus the first term in (73.6) appears
in cases where the emitting body undergoes pulsations during which its
volume changes.
Let us suppose that this is so, and determine the total intensity of the
emitted sound. The volume Aira of the fluid which flows through the closed
surface must, by the foregoing argument, be equal to the change per unit time
in the volume V of the body, i.e. to the derivative dVjdt (the volume V
being a given function of the time) : 4rra = V. Thus, at distances r such
that / <^ r <^ A, the motion of the fluid is given by the function <f> = — V(t)\$Trr.
At distances r > A, however (i.e. in the "wave region"), <f> must represent an
outgoing spherical wave, i.e. must be of the form
fttrlc)
4> =  l l. (73.7)
r
Hence we conclude at once that the emitted wave has, at all distances large
compared with /, the form
V(trlc)
<f> = ~ K ' \ (73.8)
477T
which is obtained by replacing the argument t of (tV) by t—rjc.
The velocity v = grad cf> is directed at every point along the radius
vector, and its magnitude is v = d<f>jdr. In differentiating (73.8) for distances
r > A, only the derivative of the numerator need be taken, since differentiation
of the denominator would give a term of higher order in 1/r, which we neglect.
Since W{tr\c)]dr = (l/c)F(*r/c), we obtain
v = V(tr/c)n/4iTcr, (73.9)
where n is a unit vector in the direction of r.
The intensity of the sound is given by the square of the velocity, and is
here independent of the direction of emission, i.e. the emission is isotropic.
The mean value of the total energy emitted per unit time is
I = pc jtfdf = ( P /16c7T 2 ) j (7 2 /r2)d/,
where the integration is taken over a closed surface surrounding the origin.
Taking this surface to be a sphere of radius r, and noticing that the integrand
depends only on the distance from the origin, we have finally
/ = p V 2 /4ttc. (73.10)
This is the total intensity of the emitted sound. We see that it is given by
the squared second time derivative of the volume of the body.
282 Sound §73
If the body executes harmonic pulsations of frequency to, the second time
derivative of the volume is proportional to the frequency and velocity
amplitude of the oscillations, and its mean square is proportional to the
square of the frequency for a given velocity amplitude of points on the surface
of the body. For a given amplitude of the oscillations, however, the velocity
amplitude is itself proportional to the frequency, so that the intensity of
emission is proportional to co 4 .
Let us now consider the emission of sound by a body oscillating without
change of volume. Only the second term then remains in (73.6) ; we write it
<f> = div[A(t)[r]. As in the preceding case, we conclude that the general
form of the solution at all distances r p I is <f> — div[A(£— r/c)jr]. That
this expression is in fact a solution of the wave equation is seen immediately,
since the function A(t—r/c)/r is a solution, and therefore so are its derivatives
with respect to the coordinates. Again differentiating only the numerator,
we obtain (for distances r > A)
<f> = A(trjc)nlcr. (73.11)
To calculate the velocity v = grad</>, we need again differentiate only A.
Hence we have, by the familiar rules of vector analysis for differentiation
with respect to a scalar argument,
A(/r/c)n
v = —
^grad(.I),
cr
and, substituting gTad(trjc) = (l/<:)gradr = n/c, we have finally
v = n(nA)/c2r. (73.12)
The intensity is now proportional to the squared cosine of the angle between
the direction of emission (i.e. the direction of n) and the vector A; this is
called dipole emission. The total emission is given by the integral
c 3 J r
We again take the surface of integration to be a sphere of radius r, and use
spherical coordinates with the polar axis in the direction of the vector A.
A simple integration gives finally for the total emission per unit time
/=^A2. (73.13)
The components of the vector A are linear functions of the components of
the velocity u of the body (see §11). Thus the intensity is here a quadratic
function of the second time derivatives of the velocity components.
If the body executes harmonic oscillations of frequency to, we conclude
(reasoning as in the previous case) that the intensity is proportional to co 4
§73 The emission of sound 283
for a given value of the velocity amplitude. For a given linear amplitude of
the oscillations of the body, the velocity amplitude is proportional to the
frequency, and therefore the intensity is proportional to w 6 .
In an entirely similar manner we can solve the problem of the emission of
cylindrical sound waves by a cylinder of any crosssection pulsating or
oscillating perpendicularly to its axis. We shall give here the corresponding
formulae, with a view to later applications.
Let us first consider small pulsations of a cylinder, and let S = S(t) be
its (variable) crosssectional area. At distances r from the axis of the cylinder
such that / < r 4. A, where / is the transverse dimension of the cylinder,
we have similarly to (73.8)
<f> = [S(t)l27T]\ogfr y (73.14)
where /(i) is a function of time, and the coefficient of log/r is chosen so as to
obtain the correct value for the mass flux through a coaxial cylindrical surface.
In accordance with the formula for the potential of an outgoing cylindrical
wave (the first term of formula (70.2)), we now conclude that at all distances
r > / the potential is given by
trlc
c_ r ^t)df
^ Jo V[c 2 (tt'fr2]' K ' '
As r > the leading term of this expression is the same as (73.14), and
the function f(t) in the latter equation is automatically determined (we
suppose that the derivative S(t) tends sufficiently rapidly to zero as t > — oo).
For very large values of r, on the other hand (the "wave region"), the values
oi t—t' ~ rjc are the most important in the integral (73.15). We can there
fore put, in the denominator of the integrand,
(ttfi*l<* « (2r/c)(**'r/c),
obtaining
c *"f S(t')dt'
£= Li . (73.16)
Finally, the velocity v = d<f>/dr. To effect the differentiation, it is con
venient to substitute t — t' — rjc = £:
6=  I C 1 fo^fl df
the limits of integration are then independent of r. The factor l/ v /r in
front of the integral need not be differentiated, since this would give a term
284 Sound §73
of higher order in 1/r. Differentiating under the integral sign and then
returning to the variable t', we obtain
tr/c
S(t')dt'
v =
r S(f)dt'
— . (73.17)
WW i Vi<tt')r]
The intensity is given by the product iTrrpcv 2 . It should be noticed that here,
unlike what happens for the spherical case, the intensity at any instant is
determined by the behaviour of the function S(t) at all times from — oo
to t — rjc.
Finally, for translatory oscillations of an infinite cylinder in a direction
perpendicular to its axis, the potential at distances r such that / <^ r <^ A
has the form
<f> = div(A log fr), (73.18)
where A(t) is determined by solving Laplace's equation for the flow of an
incompressible fluid past a cylinder. Hence we again conclude that, at all
distances r > /,
tr/c
r Alt' )dr
<f> = div — . (73.19)
In conclusion, we must make the following remark. We have here entirely
neglected the effect of the viscosity of the fluid, and accordingly have sup
posed that there is potential flow in the emitted wave. In reality, however, we
do not have potential flow in a fluid layer of thickness ~ \/(vJco) round the
oscillating body (see §24). Hence, if the above formulae are to be applicable,
it is necessary that the thickness of this layer should be small in comparison
with the dimension / of the body:
VWco) < /• (73.20)
This condition may not hold for small frequencies or small dimensions of
the body.
PROBLEMS
Problem 1. Determine the total intensity of sound emitted by a sphere executing small
(harmonic) translatory oscillations of frequency on, the wavelength being comparable in
magnitude with the radius R of the sphere.
Solution. We write the velocity of the sphere in the form u = u e~ i(Ot ; then <f> depends
on the time through a factor e~ ib)t also, and satisfies the equation A<f>+k 2 <f> — 0, where k = tojc.
We seek a solution in the form <j> = u • grad/(r), the origin being taken at the instantaneous
position of the centre of the sphere. For / we obtain the equation u • grad( Af+k 2 f) = 0,
whence Af+k 2 f = constant. Apart from an unimportant additive constant, we therefore
have / = Ae ikT lr. The constant A is determined from the condition 8<f>/dr = u r for r = R,
and the result is
/R\3 ikr1
d> = wre ik(r  R) i — .
r \ r I 22ikRk 2 R 2
§73 The emission of sound 285
Thus we have dipole emission. At fairly large distances from the sphere, we can neglect
unity in comparison with ikr, and <f> takes the form (73.11), the vector A being
ia>
A = ue ik(r  R) R 3 
22ikRk*R 2
Noticing that (re A) 2 = iA 2 , we obtain for the total emission, by (73.13),
2tt P , , i? 6 eo 4
= — N
3c3' ' 4 + (coR/cy
For toRfc <^ 1, this expression becomes / = irpR 6 \u \ 2 co 4 l6c 6 , a result which could also be
obtained by directly substituting in (73.13) the expression A = £R s u from §11, Problem 1.
For o>i?/c^> 1 we have 1 = 27rpci? 2 u  2 /3, corresponding to formula (73.4).
The drag force acting on the sphere is obtained by integrating over the surface of the
sphere the component of the pressure forces (p' — — p(^')r=fe) in the direction of u, and is
4tt kW3 + i(2+k 2 R*)
F = — pcoR 3 u ;
see the end of §24 concerning the meaning of a complex drag force.
Problem 2. The same as Problem 1, but for the case where the radius R of the sphere is
comparable in magnitude with V( •'/<")» whilst A^> R.
Solution. If the dimension of the body is small compared with \/( ''/<«»)> then the emitted
wave must be investigated not from the equation A ^ = 0, but from the equation of motion
of an incompressible viscous fluid. The appropriate solution of this equation for a sphere is
given by formulae (1) and (2) in §24, Problem 5. At great distances the first term in (1),
which diminishes exponentially with r, may be omitted. The second term gives the velocity
v = — i(u # grad)grad(l/r). Comparison with (73.6) shows that
A = bu = lR3[l3l(il)K3/2iK2]\i,
where #c = Ry/(u>l2v), i.e. A differs from the corresponding expression for an ideal fluid
by the factor in brackets. The result is
ttdR* / 3 9 9 9 \, ,
I = J— a>M 1+ + + + uo p.
For k ^> 1 this becomes the formula given in Problem 1 , while for ic^lwe obtain
/ = 3tt P R 2 vV\uo\ 2 I2(?,
i.e. the emission is proportional to the second, and not the fourth, power of the frequency.
Problem 3. Determine the intensity of sound emitted by a sphere executing small (har
monic) pulsations of arbitrary frequency.
Solution. We seek a solution of the form <f> = {aujr)e ik ^ r ~ Ki , R being the equilibrium
radius of the sphere, and determine the constant a from the condition [d4>Jdr] r =R = u
= u e~ i0)t (where u is the radial velocity of points on the surface of the sphere):
a = R 2 l{ikR\).
The intensity is I = 27rpcu  2 # ! R 4 /(l +k 2 R i ). For kR < 1, / = 2ir P a) 2 R i \u \ 2 /c, in accor
dance with (73.10), while for kR ^> 1, / = 2npcR 2 \u \ 2 , in accordance with (73.4).
Problem 4. Determine the nature of the wave emitted by a sphere (of radius R) executing
small pulsations, when the radial velocity of points on the surface is an arbitrary function
u(t) of the time.
286 Sound §73
Solution. We seek a solution in the form <f> =f(t')Jr, where t' = t—{r—R)jc, and deter
mine / from the boundary condition d<f>j8r = u(t) for r = R. This gives the equation
dfJdt+cf(t)JR = — Rcu(t). Solving this linear equation and replacing t by t' in the solution,
we obtain
„ t
cR r
<f,(r, t') = e<*/R u{t)(*»/R&t. (1)
—oo
If the oscillations of the sphere cease at some instant, say t — (i.e. u(j) = for r > 0),
then the potential at a distance r from the centre will be of the form <f> = constant X e~ e ' ls
after the instant t = (r—R)Jc, i.e. it will diminish exponentially.
Let T be the time during which the velocity u(t) changes appreciably. If T ^> Rjc, i.e. if
the wavelength of the emitted waves A ~ cT^> R, then we can take the slowly varying factor
«(t) outside the integral in (1), replacing it by u(t'). For distances r^> R, we then obtain
<f> = ~(R 2 Jr)u(t—rjc), in accordance with formula (73.8). If, on the other hand, T <^.RJc,
we obtain in a similar manner
cR C
$ = w ( T )d T , v = 8^1 8r = {Rlr)u{t'),
—oo
in accordance with formula (73.4).
Problem 5. Determine the motion of an ideal compressible fluid when a sphere of radius
R executes in it an arbitrary translatory motion, with velocity small compared with that of
sound.
Solution. We seek a solution in the form
$ = div[f(*>], (1)
where r is the distance from the origin, taken at the position of the centre of the sphere at
the time t' = t—(r—R)/c; since the velocity u of the sphere is small compared with the
velocity of sound, the movement of the origin may be neglected. The fluid velocity is
„ aa 3(f.n)nf t 3(f'.n)nf' , (f".n)n
v = grad^ = + + , (2)
where n is a unit vector in the direction of r, and the prime denotes differentiation with
respect to the argument of f. The boundary condition is v r = u • n for r — R, whence
f"(t)+(2clR)f'(t)+(2c*IR 2 )f(t) = Rc 2 u(t). Solving this equation by variation of the para
meters, we obtain for the function f(t) the general expression
* ( \
f(t) = cR*e*i* fu(T)sin^— IV/Rc1t. (3)
—00
In substituting in (1), we must replace t by t'. The lower limit is taken as — oo so thatf
shall be zero for t = — oo.
Problem 6. A sphere of radius i? begins at time t = to move with constant velocity u .
Determine the sound intensity emitted at the instant when the motion begins.
Solution. Putting in formula (3) of Problem 5 u(t) = f or t < and u(t) = u for
t > 0, and substituting in formula (2) (retaining only the last term, which decreases least
rapidly with r), we find the fluid velocity far from the sphere :
v = — n(n«uo)
\/2R l ct' \
er<*i R sin I M,
§73 The emission of sound 287
where t' > 0. The total intensity diminishes with time according to
I = (87rl3)cpR 2 uo 2 e^' / Rsm z (ct'IRi7r).
The total amount of energy emitted is iTrpi? 3 u 2 .
Problem 7. Determine the intensity of sound emitted by an infinite cylinder, of radius R,
executing harmonic pulsations of wavelength A ^> R.
Solution. According to formula (73.14), we find first of all that, at distances r <^ A
(in Problems 7 and 8 r is the distance from the axis of the cylinder), the potential is
<f> = Ru log kr, where u = u a e~ im is the velocity of points on the surface of the cylinder.
From a comparison with formulae (70.7) and (70.8), we now find that at large distances the
potential is of the form 4> = —Ru^{iirl2kr)e ikr . The velocity is therefore
v = Ru^/(7rk/2ir)ne ikr ,
where n is a unit vector perpendicular to the axis of the cylinder, and the intensity per unit
length of the cylinder is / = $n 2 po>R 2 u 2 .
Problem 8. Determine the intensity of sound emitted by a cylinder executing harmonic
translatory oscillations in a direction perpendicular to its axis.
Solution. At distances r <^ A we have <j> — — div(i? 2 u log kr) ; cf. formula (73.18) and
§10, Problem 3. Hence we conclude that at large distances
<j> = ^V^^divC^^u/Vr) = R 2 u.n^/(7Tkl2ir)eMr t
whence the velocity is v = — kR 2 \/(inkl2r)n(vfn)e ikr . The intensity is proportional to the
squared cosine of the angle between the directions of oscillation and emission. The total
intensity is J = (7r 2 /4c 2 )pa) 3 i2 4 u  2 .
Problem 9. Determine the intensity of sound emitted by a plane surface whose temperature
varies periodically with frequency o» <^ e 2 /x» where x is the thermometric conductivity of the
fluid.
Solution. Let the variable part of the temperature of the surface be T / 9 e~ i<ot . These
temperature oscillations cause a damped thermal wave in the fluid (52.17):
T' = T , oe i<ot e~ a  i) ^ <  b>/2 x )x i
and the fluid density therefore oscillates also: />' = (dp\dT) v T = —pPT', where P is the
coefficient of thermal expansion. This, in turn, results in the occurrence of a motion deter
mined by the equation of continuity: p dv/dx = —dp'jdt = —imp^T'. At the solid surface
the velocity v x = v = 0, and far from the surface it tends to the limit
00 ^
v = imp f T'dx = ^Lp^ajfiT'oer**
o
This value is reached at distances , ^ J V(xl <0 )> which are small compared with c/to, and we
thus have a boundary condition on the resulting sound wave. Hence we find the intensity
per unit area of the surface to be / = icp^ 2 wx\T / e \ 2 .
Problem 10. A point source emitting a spherical wave is at a distance / from a solid wall
which totally reflects sound and bounds a halfspace occupied by fluid. Determine the ratio
of the total intensity of sound emitted by the source to that which would be found in an infinite
medium, and the dependence of the intensity on direction for large distances from the source.
Solution. The sum of the direct and reflected waves is given by a solution of the wave
equation such that the normal velocity component v n — 3<f>/dn is zero at the wall. Such a
solution is
oikr pikr'
 + —)<»
288 Sound §74
(we omit the constant factor, for brevity), where r is the distance from the source O (Fig.
39), and r' is the distance from a point O' which is the image of O in the wall. At large dis
tances from the source we have r' 7H r— 21 cos 0, so that
e i(kr— <ot)
(f) = (lf e 2ta?cos<?).
r
The dependence of the intensity on direction is given by a factor cos 2 (&/ cos 0).
To determine the total intensity, we integrate the energy flux q = p'v = — p<£ grad 4
(see (64.4)) over the surface of a sphere of arbitrarily small radius, centred at O. This gives
2irpk<*>{\ +[l/2&/] sin 2kl). In an infinite medium, on the other hand, we should have simply a
spherical wave <p = e i(*r«t)/ r> w ith a total energy flux 2irpka>. Thus the required ratio of
intensities is 1 +(l/2kl) sin 2kl.
3 6 e  / 
^/^.
Fig. 39
Problem 11. The same as Problem 10, but for a fluid bounded by a free surface.
Solution. At the free surface the condition p' = — p4> = must hold; in a monochro
matic wave this is equivalent to ^ = 0. The corresponding solution of the wave equation is
oikr t>ikf
\g—io>t
(gtKr e i/cr \
At large distances from the source, the intensity is given by a factor sin 2 (kl cos 8). The re
quired ratio of intensities is 1 —(\j2kl) sin 2kl.
§74. The reciprocity principle
In deriving the equations of a sound wave in §63, it was assumed that the
wave is propagated in a homogeneous medium. In particular, the density
po of the medium and the velocity of sound in it, c, were regarded as constants.
In order to obtain some general relations applicable for an arbitrary inhomo
geneous medium, we shall first derive the equation for the propagation of
sound in such a medium.
We write the equation of continuity in the form dpjdt+pdivv = 0.
Since the propagation of sound is adiabatic, we have
dt \8plsdt c* dt c*\dt * F ]
and the equation of continuity becomes #p/d* + v«grad/> + pc 2 div v = 0.
As usual, we put p — po + p', where po is now a given function of the
coordinates. In the equation p = po +p', however, we must put as before
§74 The reciprocity principle 289
po = constant, since the pressure must be constant throughout a medium in
equilibrium (in the absence of an external field, of course). Thus we have to
within secondorder quantities dp'/dt+poc 2 div v = 0.
This equation is the same in form as equation (63.5), but the coefficient
poc 2 is a function of the coordinates. As in §63, we obtain Euler's equation
in the form dvjdt = (1/po) gradp'. Eliminating v, and omitting the
suffix in po, we finally obtain the equation of propagation of sound in an
inhomogeneous medium:
&&L12L0. (74,)
p pc l ct*
If the wave is monochromatic, with frequency w, we have p' = — co 2 />',
so that
d . v grad/ + ^, = o (742)
P P&
Let us consider a sound wave emitted by a pulsating source of small
dimension; we have seen in §73 that the emission is isotropic. We denote by
A the point where the source is, and by pjfi) the pressure p' at a point B
in the emitted wave.f If the same source is placed at B, it produces at A
a pressure which we denote by Pb(A). We shall derive the relation between
p A (B) and p B (A).
To do so, we use equation (74.2), applying it first to the sound from a
source at A and then to the sound from a source at B:
eradp' A (o 2 , n ,. gradp' B co 2
d iv« L± + p' A = 0, div2 — + —Tp's = 0.
P pc 2 p pr
We multiply the first equation by p '# and the second by p' a and subtract.
The result is
, gra&pA . ,. gradp's
p' B div p a div
P P
_ di / P'b g™&P'A P'a gradp'e \ = Q
\ P P I
We integrate this equation over the volume between an infinitely distant
closed surface C and two small spheres Ca and Cb which enclose the points
A and B respectively. The volume integral can be transformed into three
surface integrals, and the integral over C is zero, since the sound field vanishes
at infinity. Thus we obtain
r ^ *«£*__, VWb\ m _ i (74 . 3)
t The dimension of the source must be small compared with the distance between A and B and
with the wavelength.
290 Sound §74
Inside the small sphere C A , the pressure p' A in the wave from a source
at A falls off rapidly with the distance from A, and the gradient gradp' A
is therefore large. The pressure p' B due to a source at B is a slowly varying
function of the coordinates in the region near the point A, which is at a
considerable distance from B, so that the gradient gradp' B is relatively small.
When the radius of the sphere C A is sufficiently small, therefore, we can
neglect the integral
j(p'A/p)gradp' B d£
over C A in comparison with
J(P'b/p) gradp' A .df,
and in the latter the almost constant quantity p' B can be taken outside the
integral and replaced by its value at the point A. Similar arguments hold for
the integrals over the sphere C B , and as a result we obtain from (74.3) the
relation
But (l[p)gradp' = dvfdt, and this equation can therefore be rewritten
a
c A
The integral
p' B (A)—j>v A .df = p' A (B)jj>v B .df.
<fv^.df
over C A is the volume of fluid flowing per unit time through the surface of
the sphere C A , i.e. it is the rate of change of the volume of the pulsating
source of sound. Since the sources at A and B are identical, it is clear that
fv^.df=fv B .df,
and consequently
p' A {B) = p' B {A). (74.4)
This equation constitutes the reciprocity principle : the pressure at B due
to a source at A is equal to the pressure at A due to a similar source at B.
It should be emphasised that this result holds, in particular, for the case
where the medium is composed of several different regions, each of which
§75 Propagation of sound in a tube 291
is homogeneous. When sound is propagated in such a medium, it is reflected
and refracted at the surfaces separating the various regions. Thus the reci
procity principle is valid also in cases where the wave undergoes reflection
and refraction on its path from A to B.
PROBLEM
Derive the reciprocity principle for dipole emission of sound by a source which oscillates
without change of volume.
Solution. In this case the integral
^V^df
over C A is zero identically, and the next approximation must be taken in calculating the
integrals in (74.3). To do so, we write, as far as the firstorder terms,
P'b = p'^ + Tffradp's,
where r is the radius vector from A. In the integral
, grad^ , gr ade's \ Ar (1)
f^*^,j£±t>yn
the two terms are now of the same order of magnitude. Substituting here for p' B the above
expansion, and using the fact that the integral
j>(ll P )gradp' A >d£
over Ca is now zero, we obtain
<b (r grade's) p a j
Next, we take the almost constant quantity gradp' B = pv B outside the integral, replacing
it by its value at A :
C A P
where pa is the density of the medium at the point A To calculate this integral, we notice
that near a source the fluid can be supposed incompressible (see §73), and hence we can write
for the pressure inside the small sphere Ca, by (11.1), P'a = p+ = pAr/r 3 . In a mono
chromatic wave v = — icov, A = —itoA; introducing also the unit vector tla in the direction
of the vector A for a source at A, we find that the integral (1) is proportional to p A v B (A) • n^.
Similarly, the integral over the sphere C B is proportional to — pbva(B) • n Bl with the same
factor of proportionality. Equating the sum to zero, we find the required relation
PA^B(A)'n A = pBVA{B)n B ,
which expresses the reciprocity principle for dipole emission of sound.
§75. Propagation of sound in a tube
Let us now consider the propagation of a sound wave in a long narrow tube.
By a "narrow" tube we mean one whose width is small compared with the
292 Sound
§75
wavelength. The crosssection of the tube may vary along its length in both
shape and area. It is important, however, that this variation should occur
fairly slowly: the crosssectional area S must vary only slightly over distances
of the order of the width of the tube.
Under these conditions we can suppose that all quantities (velocity,
density, etc.) are constant over any transverse crosssection of the tube. The
direction of propagation of the wave can be supposed to coincide with that of
the axis of the tube at all points. The equation for the propagation of such
a wave is most conveniently derived by a method similar to that used in §13
in deriving the equation for the propagation of gravity waves in channels.
In unit time a mass Spv of fluid passes through a crosssection of the tube.
Hence the mass of fluid in the volume between two transverse crosssections
at a distance dx apart decreases in unit time by
{S P v) x+&x {Spv) x = [d(S P v)/8x]dx,
the coordinate x being measured along the axis of the tube. Since the volume
between the two crosssections remains constant, the decrease must be due
only to the change in density of the fluid. The change in density per unit time
is dp/dt, and the corresponding decrease in the mass of fluid in the volume
S dx between the two crosssections is S{8pj8t)dx. Equating the two
expressions, we obtain
S8p/8t = 8(Spv)ldx, (75.1)
which is the "equation of continuity" for flow in a pipe.
Next, we write down Euler's equation, omitting the term quadratic in the
velocity:
dv/dt= (l/p)8p/8x. (75.2)
We differentiate (75.1) with respect to time, regarding p on the righthand
side as independent of time, since the differentiation of p gives a term which
involves v dp/dt = v dpjdt and is therefore of the second order of smallness.
Thus S 8 2 p\8t 2 =  8{Sp8vj8t)J8x. Here we substitute the expression (75.2)
for 8v/8t, and express the derivative of the density on the lefthand side in
terms of the derivative of the pressure by p = (8p[8p)p = p/c 2 .
The result is the following equation for the propagation of sound in a
tube:
1 8 I dp\ 1 8 2 p
In a monochromatic wave p depends on time through a factor e _iw ', and
(75.3) becomes
1 8 l 8p\
sT x \ S £) +k2p = (i  < 75 ' 4 >
where k — co/c is the wave number.f
f Here, and in the Problems, p denotes the variable part of the pressure, which we have previ
ously denoted by p'.
§75 Propagation of sound in a tube 293
Finally, let us consider the problem of the emission of sound from the
open end of a tube. The pressure difference between the gas in the end of
the tube and that in the space surrounding the tube is small compared with the
pressure differences within the tube. Hence the boundary condition at the
open end of the tube is, with sufficient accuracy, that the pressure p should
vanish. The gas velocity v at the end of the tube is not zero ; let its value be
vq. The product Svq is the volume of gas leaving the tube per unit time.
We can now regard the open end of the tube as a "source" of gas of strength
Svq. The problem of the emission from a tube thus becomes equivalent to
that of the emission by a pulsating body, which is solved by formula (73.10).
In place of the time derivative Voi the volume of the body we must now put
Svo. Thus the total intensity of the sound emitted is
/ = pS^IAttc. (75.5)
PROBLEMS
Problem 1. Determine the transmission coefficient for sound passing from a tube of cross
section S x into one of crosssection S 2 .
Solution. In the first tube we have two waves, the incident wave p x — a 1 e iikx ~ a>t) and
the reflected wave p x ' = a{ ' e M*+<»t) . In the second tube we have the transmitted wave
p 2 = a 2 e^ kx  m \ At the point where the tubes join (x = 0), the pressures must be equal,
and so must the volumes Sv of gas passing from one tube to the other per unit time. These
conditions give aj+fli" = <h, S 1 (a 1 a 1 ') = S 2 a 2 , whence a 2 = 2a 1 5 1 /( ( S , 1 + 4 S 2 ). The ratio D
of the energy flux in the transmitted wave to that in the incident wave is
D = S 2 pc\^\zjS iP c\^\z = S 2 ^/£iM2
or
4SlS2 , /S2Si\ 2
D =
\s*+sj
Problem 2. Determine the amount of energy emitted from the open end of a cylindrical
tube.
Solution. In the boundary condition p = at the open end of the tube, we can approxi
mately neglect the emitted wave (we shall see that the intensity emitted from the end of the
tube is small). Then we have the condition p x = —p x \ where p t andpi are the pressures in
the incident wave and in the wave reflected back into the tube ; for the velocities we have
correspondingly v t = Vy, so that the total velocity at the end of the tube is v = Vx+v^ = 2v t .
The energy flux in the incident wave is cSpv} = icSpv^. Using (75.5), we obtain for the
ratio of the emitted energy to the energy flux in the incident wave D = Su) 2 /nc 2 . For a
tube of circular crosssection (radius R) we have D = i? a w 2 /c 2 . Since, by hypothesis R<€clo>
it follows that D<1. » \ / ,
Problem 3. One of the ends of a cylindrical pipe is covered by a membrane which executes
a given oscillation and emits sound ; the other end is open. Determine the way in which sound
is emitted from the tube.
Solution. In the general solution
p = (a e ikx + J) e i1cxyiut
we determine the constants a and b from the conditions v = u = M e _i& ", the given velocity
of the membrane, at the closed end (x = 0), and p = at the open end (x = /). These give
294 Sound §76
ae lke +be~ m = 0, a—b = cpu . Determining a and b, we find the gas velocity at the open
end of the tube to be v — w/cos kl. If the tube were absent, the intensity of the sound emitted
by the oscillating membrane would be given by the mean square <S 2 Itt 2 = 5 2 w 2 m 2 , according
to formula (73.10) with Su in place of V; S is the crosssectional area of the membrane.
The emission from the end of the tube is proportional to S 2 \v \ 2 a) 2 . Defining the "amplifi
cation coefficient" of the pipe as A = S a \v \*/S a \u[*, we obtain A — 1/cos 2 kl. This becomes
infinite for frequencies of oscillation of the membrane equal to the characteristic frequencies
of the tube {resonance) ; in reality, of course, it remains finite because of effects which we have
neglected (such as friction due to the emission of sound).
Problem 4. The same as Problem 3, but for a conical tube, with the membrane covering
the smaller end.
Solution. The crosssection of the tube is S = SqX 2 ; let the values of the coordinate x
which correspond to the smaller and larger ends be x lt x 2 , so that the length of the tube is
/ = x 2 —Xi. The general solution of equation (75.4) is p = (l/x)(ae ilcx +be~ ikx )e~ im ; a and b
are determined from the conditions v = u for x = Xi and p = for * = x 2 . The amplifica
tion coefficient is found to be
SoX2 2 \v2\ 2 k 2 Xi 2
A
Sox± 2 1 u  2 (sin kl + kxi cos kl) 2
Problem 5. The same as Problem 3, but for a tube whose crosssection varies exponen
tially along its length: S = Soe xx .
Solution. Equation (75.4) becomes d 2 pl8x 2 +adpjdx+k 2 p = 0, whence
p = e \*x( ae imx + b e imxyiuit t
with m = \/{k 2 —\a. 2 ). Determining a and b from the conditions v = u for x = and p —
for x = /, we find the amplification coefficient
Soe al \vq\ 2 e al
So\u\ 2 [£( a / m ) s i n m ^+ cos m ^] 2
for k > $a and
[(a/m') sinh m'l+ cosh m'Vf V KZ h
for k < ^a.
§76. Scattering of sound
If there is some body in the path of propagation of a sound wave, then the
sound is scattered: besides the incident wave there appear other (scattered)
waves, which are propagated in all directions from the scattering body. The
scattering of a sound wave occurs simply on account of the presence of the
body in its path. In addition, the incident wave causes the body itself to move,
and this in turn brings about additional emission of sound by the body, i.e.
further scattering. If, however, the density of the body is large compared
with that of the medium in which the sound is propagated, and its compres
sibility is small, then the scattering due to the motion of the body forms only a
small correction to the main scattering caused by the mere presence of the
body. In what follows we shall neglect this correction, and therefore suppose
the scattering body immovable.
§76 Scattering of sound 295
We assume that the wavelength A of the sound is large compared with the
dimension / of the body; to calculate the properties of the scattered wave,
we can then use formulae (73.8) and (73.11).j In doing so, we regard the
scattered wave as being emitted by the body; the only difference is that,
instead of a motion of the body in the fluid, we now have a motion of the fluid
relative to the body. The two problems are clearly equivalent.
For the potential of the emitted wave we have obtained the expression
<f> = — VjAttt — A* rjcr 2 . In this formula V was the volume of the body.
In the present case, however, the volume of the body itself remains unchanged,
and V must be taken not as the rate of change of the volume of the body, but
as the volume of fluid which would enter, per unit time, the volume Vq
occupied by the body if the body were absent. For, in the presence of the
body, this volume of fluid does not penetrate into Vq, which is equivalent to
the emission of the same volume of fluid from Vq. The coefficient of 1/477T
in the first term of <f> must, as we have seen in §73, be just the volume of fluid
emitted from the origin per unit time. This volume is easily found. The
change per unit time in the mass of fluid in a volume equal to that of the body
is Vop, where p gives the rate of change of the fluid density in the incident
sound wave (since the wavelength is large compared with the dimension
of the body, the density p may be supposed constant over distances of the
order of this dimension; hence we can write the rate of change of the mass of
fluid in Vq as Vop simply, where p is the same throughout the volume Vq).
The change in volume corresponding to a mass change Vop is evidently
Vopfp. Thus V in the expression for (f> must be replaced by Vop/ p. In an
incident plane wave, the variable part p of the density is related to the velocity
by p = pv/c; hence p = p' = pvfc, and we can replace Vopfp by Vqvjc.
When the body moves in the fluid, the vector A is determined by formulae
(11.5), (11.6): AnpAi = miicUk+ pVoUi. We must now replace the velocity
u of the body by the reversed velocity v of the fluid in the incident wave which
it would have at the position of the body if the latter were absent. Thus
At = —m ilc V]clATTp—VoViJ^TT. (76.1)
We finally obtain for the potential of the scattered wave
0sc= Vov\\Trcrk*T\cr\ (76.2)
the vector A being given by formula (76.1). Hence we have for the velocity
distribution in the scattered wave
v sc = Vq vnl47rrc 2 + n(n • A)frc 2 (76.3)
(see §73), n being a unit vector in the direction of scattering.
The mean amount of energy scattered per unit time into a given solid angle
element do is given by the energy flux, which is cpv S c 2 r 2 do. The total scat
tered intensity 7 S c is obtained by integrating this expression over all directions.
f At the same time, the dimension of the body must be large in comparison with the displacement
amplitude of fluid particles in the wave, since otherwise the fluid is not in general in potential flow.
296 Sound §76
The integration of twice the product of the two terms in (76.3) gives zero,
since this product is proportional to the cosine of the angle between the
direction of scattering and the direction of propagation of the incident wave,
and there remains (cf. (73.10) and (73.13))
The scattering is generally characterised by what is called the effective
crosssection do, which is the ratio of the (time) average energy scattered into
a given solidangle element to the mean energy flux density in the incident
wave. The total effective crosssection a is the integral of da over all directions
of scattering, i.e. it is the ratio of the total scattered intensity to the incident
energy flux density, and evidently has the dimensions of area.
The mean energy flux density in the incident wave is cp\ 2 . Hence the
differential effective scattering crosssection is (cpv S c 2 /c/>v 2 )r 2 do, i.e.
do = (v^2/v2)r2do. (76.5)
The total effective crosssection is
V 2 V 2 " 4tt A*
a = t= + — •=■. (76.6)
477c 4 v2 3c 4 V2 v '
For a monochromatic incident wave, the mean square second time derivative
of the velocity is proportional to the fourth power of the frequency. Thus the
effective crosssection for the scattering of sound by a body which is small
compared with the wavelength is proportional to oA.
Finally, let us briefly discuss the opposite limiting case, where the wave
length of the scattered sound is small compared with the dimension of the
body. In this case all the scattering, except for the scattering through
very small angles, amounts to simple reflection from the surface of the body.
The corresponding part of the total effective scattering crosssection is clearly
equal to the area S of the crosssection of the body by a plane perpendicular
to the direction of the incident wave. The scattering through small angles
(of the order of A//), however, constitutes diffraction from the edges of the
body. We shall not pause here to expound the theory of this phenomenon,
which is entirely analogous to that of the diffraction of light.f We shall only
mention that, by Babinet's principle, the total intensity of diffracted sound
is equal to the total intensity of reflected sound. Hence the diffraction
part of the effective scattering crosssection is also equal to S, and the total
crosssection is therefore 2S.
PROBLEMS
Problem 1 . Determine the effective crosssection for the scattering of a plane sound wave
by a solid sphere of radius R small compared with the wavelength.
t See The Classical Theory of Fields, §§77 to 79.
§76 Scattering of sound 297
Solution. The velocity at a given point in a plane wave is v = a cos wt. In the case of a
sphere (see §11, Problem 1), the vector A is — £R 3 v. For the differential effective crosssection
we obtain
da = (1f COS 0)2 do,
9c 4
where 8 is the angle between the direction of the incident wave and the direction of scattering.
The scattered intensity is greatest in the direction 8 = n, which is opposite to the direction
of incidence. The total effective crosssection is
a = (77r/9)( J R3 G> 2/ c 2)2. (1)
Here (and also in Problems 3 and 4 below) it is assumed that the density p of the sphere
is large compared with the density p of the gas ; if this were not so, it would be necessary
to take account of the movement of the sphere by the pressure forces exerted on it by the
oscillating gas.
Problem 2. Determine the effective crosssection for the scattering of sound by a drop of
fluid, taking into account the compressibility of the fluid and the motion of the drop caused
by the incident wave.
Solution. When the pressure of the gas surrounding the drop changes adiabatically by />',
the volume of the drop is reduced by (V lp )(dp<>lc>p) s p', where p is the density of the drop.
But (dp/dpo), is the square of the velocity of sound c in the fluid, and the pressure change in a
plane wave is related to the velocity by p' = vcp, where p is the density of the gas. Thus
the decrease in the volume of the drop is V vcpjc 2 p per unit time. In the expressions (76.2)
and (76.3), we must now replace V vjc by the difference V {vlc—vcpjc 2 p^). Moreover, in
the expression for A we must replace — v by the difference u— v, where u is the velocity
acquired by the drop as a result of the action of the incident wave. For a sphere we have, using
the results of §11, Problem 1, A = /? 3 v(P _ Po)/(2po + P) Substituting these expressions, we
have the effective crosssection
co*R«l/ c*p \ pop )*
do = {II — 3 cos 9
9c4 \\ c 2p / 2po+p
The total effective crosssection is
4rrftAR6 // <fip \2 3(/> />) 2
I \ Cn 2 po /
+
9c* \\ coW (2p +p)2)
Problem 3. Determine the effective crosssection for the scattering of sound by a solid
sphere of radius R small compared with \/(vjw). The specific heat of the sphere is supposed
so large that its temperature can be regarded as a constant.
Solution. In this case we have to take into account the effect of the gas viscosity on the
motion of the sphere, and the vector A must be modified as shown in §73, Problem 2. For
RV(<oJv) < 1 we have A = —3iRvvJ2o).
The thermal conductivity of the gas also results in scattering of the same order. Let
TV  ' 6 " be the temperature variation at a given point in the sound wave. The temperature
distribution near a sphere is (see §52, Problem 2)
(for r = R we must have T' = 0). The amount of heat transferred from the gas to the sphere
per unit time is (for RVHx) < 1) 9 = 4irR z k[&T I&r\ r=R = 4nR K T &**>*. This transfer
of heat results in a change in the volume of the gas, which can be taken to affect the scattering
like a corresponding effective change in the volume of the sphere, V — — AirRxfiT' Q e~ im
= — AirRx{y — l)v/c, where ]8 is the coefficient of thermal expansion of the gas and y = c p /c v ;
we have used also formulae (63.13) and (77.2).
298 Sound §77
Taking account of both effects, we obtain the differential effective scattering crosssection
da = (a)Rlc*)*\x(y  1)  \v cos df do.
The total effective crosssection is
a = 477(ft> J R/c 2 ) 2 [x 2 (yl)2 + v2].
These formulae are valid only if the Stokes frictional force is small compared with the
inertia force, i.e. tjR <^ Mco, where M = 4irR 3 p /3 is the mass of the sphere ; otherwise, the
movement of the sphere by viscosity forces becomes important.
Problem 4. Determine the mean force on a solid sphere which scatters a plane sound wave
(A>#).
Solution. The momentum transmitted per unit time from the incident wave to the sphere,
i .e. the required force, is the difference between the momentum in the incident wave and the
total momentum flux in the scattered wave. From the incident wave an energy flux acE
is scattered, where E is the energy density in the incident wave ; the corresponding momen
tum flux is obtained by dividing by c, and is therefore oE . In the scattered wave, the momen
tum flux into the solid angle element do is E sc r 2 do = E da; projecting this on the direction
of propagation of the incident wave (which is obviously the direction of the required force),
and integrating over all angles, we obtain
Eq COS 6 da.
Thus the force on the sphere is
F = Eo J(lcos0)da.
Substituting for da from Problem 1, we obtain F = UircoWE^fic*.
§77. Absorption of sound
The existence of viscosity and thermal conductivity results in the dissipa
tion of energy in sound waves, and the sound is consequently absorbed,
i.e. its intensity progressively diminishes. To calculate the rate of energy
dissipation i£mech> we use the following general arguments. The mechanical
energy is just the maximum amount of work that can be done in passing
from a given nonequilibrium state to one of thermodynamic equilibrium.
As we know from thermodynamics,^ the maximum work is obtained when the
transition is reversible (i.e. without change of entropy), and is then
#mech = Bq — E(S), where Eo is the given initial value of the energy, and
E(S) is the energy in the equilibrium state with the same entropy S as the
system had initially. Differentiating with respect to time, we obtain
$mech = —£(S) = (dEIdS)S. The derivative of the energy with respect
to the entropy is the temperature. Hence dE/dS is the temperature which
the system would have if it were in thermodynamic equilibrium (with the
given value of the entropy). Denoting this temperature by To, we therefore
have .Cmech — — TqS.
t See, for instance, Statistical Physics, §19.
§77 Absorption of sound 299
We use for 3 the expression (49.6), which gives the rate of change of the
entropy due to both thermal conduction and viscosity. Since the temperature
T varies only slightly through the fluid, and differs little from To, it can be
taken outside the integral, and To can be written as T simply:
k r f / OVi uVk OVi \ *
&  Y /(«n«ID.dF i ,J( + » ^) ir
^J(divv)2dr. (77.1)
This formula generalises formula (16.3) to the case of a compressible fluid
which conducts heat.
Let the #axis be in the direction of propagation of the sound wave. Then
v x = v cos(kxwt), v y = v z = 0. The last two terms in (77.1) give
to + Q\(^\W = Wto + frx? jsm*(kxa>t)dV.
We are, of course, interested only in the time average; taking this average,
we have & 2 (ji?+ £) . ^o 2 ^o, where Vq is the volume of the fluid.
Next we calculate the first term in (77.1). The deviation V of the tem
perature in the sound wave from its equilibrium value is related to the
velocity by formula (63.13), so that the temperature gradient is
dTjdx = {^cTjc p )dvjdx = (pcTlc p )voksm(kxcot).
For the time average of the first term in (77. 1) we obtain  KC 2 T(Pvo 2 k 2 Vol2c p 2 .
Using the wellknown thermodynamic formulae
c p c v = W(dp\d P ) T = T^{c v \c v \dp\d P ) s = TpWc v /c p , (77.2)
we can rewrite this expression as — \k(\jc v — \jc p )k 2 vo 2 Vo.
Collecting the above results, we find the mean value of the energy dissi
pation :
imech= lVo[(^ + + K(l/^iy. (77.3)
The total energy of the sound wave is
E = \pvoW . (77.4)
The damping coefficient derived in §25 for gravity waves gives the manner
of decrease of the intensity with time. For sound, however, the problem is
usually stated somewhat differently: a sound wave is propagated through a
fluid, and its intensity decreases with the distance x traversed. It is evident
that this decrease will occur according to a law e~ 2yx > and the amplitude will
decrease as e~ yx , where the absorption coefficient y is defined by
y = l^mechl/2^. (77.5)
300 Sound §77
Substituting here (77.3) and (77.4), we find the following expression for the
sound absorption coefficient:
r^lto+o+iii)] (77  6)
We may point out that it is proportional to the square of the frequency of
the sound.f
This formula is applicable so long as the absorption coefficient determined
by it is small: the amplitude must decrease relatively little over distances of
the order of a wavelength (i.e. we must have ycjco <^ 1). The above deriva
tion is essentially founded on this assumption, since we have calculated
the energy dissipation by using the expression for an undamped sound wave.
For gases this condition is in practice always satisfied. Let us consider, for
example, the first term in (77.6). The condition ycjco <^ 1 means that
vcojc 2 <^ 1. It is known from the kinetic theory of gases, however, that the
viscosity coefficient v for a gas is of the order of the product of the mean
free path / and the mean thermal velocity of the molecules; the latter is of
the same order as the velocity of sound in the gas, so that v ~ Ic. Hence we
have
vco/c 2 ~ lco/c ~ //A <^ 1, (77.7)
since we know that / <^ A. The thermalconduction term in (77.6) gives the
same result, since x ~ v 
In liquids, the condition of small absorption is always fulfilled when the
problem of sound absorption, as stated here, is significant at all. The absorp
tion over one wavelength can become large only if the viscosity forces
are comparable with the pressure forces which occur when the substance is
compressed. In these conditions, however, the NavierStokes equation itself
(with the viscosity coefficients independent of frequency) becomes invalid
and a considerable dispersion of sound, due to processes of internal friction,
occurs. %
For absorption of sound, the relation between the wave number and the
frequency can evidently be written
k = co/c+iaco 2 , (77.8)
where a denotes the coefficient of a> 2 in the absorption coefficient y = aco 2 .
t M. A. Isakovich has shown that there must be a special absorption when sound is propagated
in a twophase system (an emulsion). Because of the different thermodynamic properties of the two
components, their temperature changes during the passage of the sound wave will in general be
different. The resulting heat exchange between the components leads to an additional absorption of
sound. On account of the relative slowness of this heat exchange, a considerable dispersion of the
sound takes place comparatively quickly. For detailed calculations see M. A. Isakovich, Zhurnal
experimental' noi i teoreticheskoi fiziki 18, 907, 1948.
J A special case where strong absorption is possible but can be discussed by the usual methods is
that of a gas with a thermal conductivity which is unusually large compared with its viscosity, on
account of effects such as radiative transfer at very high temperatures (see Problem 3).
§77 Absorption of sound 301
It is easy to see from this how the equation for a travelling sound wave must
be modified in order to take absorption into account. To do so, we notice that,
in the absence of absorption, the differential equation for (say) the pressure
p' = p'{xct) can be written dp' fix = {\jc)dp'jdt. The equation whose
solution is e i(kx ~ wt \ with k given by (77.8), must clearly be
v = _r_v + a ?v (77 . 9)
dx c 8t dt 2
If we replace t by t + x/c, this equation becomes
dp' fix = adty'ldT*,
i.e. a onedimensional equation of thermal conduction.
The general solution of this equation can be written (see §51)
p'(x,r) = I f/o(r')exp[(r'T)2/te]dT', (77.10)
where p'o(r) = p'(0, r). If the wave is emitted during a finite time interval,
this expression becomes, at sufficiently large distances from the source,
p'(x, r) = „ I exp(  r2/te) f p'o{r') dr'. (77.1 1)
2y(7rax) J
In other words, the wave profile at large distances is Gaussian. Its "width"
is of the order of <\/(ax), i.e. it increases as the square root of the distance
travelled by the wave, while the amplitude falls off inversely as y/x. Hence
we at once conclude that the total energy of the wave decreases as l/y/x.
It is easy to derive analogous formulae for spherical waves; to do so, we
must use the fact that for such a wave
jp'dt =
(see §69). Instead of (77.11) we now have
1 d exp(T 2 /4«r)
p'(r, t) = constant x
r dr \/r
or
T
p'(r,r) = constant x— exp(r 2 /4ar). (77.12)
fZ
Strong absorption must occur when a sound wave is reflected from a solid
wall (K. F. Herzfeld, 1938; B. P. Konstantinov, 1939). The reason for
this is the following. In a sound wave not only the density and the pressure,
but also the temperature, undergo periodic oscillations about their mean values.
Near a solid wall, therefore, there is a periodically fluctuating temperature
difference between the fluid and the wall, even if the mean fluid temperature is
302 Sound §77
equal to the wall temperature. At the wall itself, however, the temperatures
of the wall and the adjoining fluid must be the same. As a result, a large
temperature gradient is formed in a thin boundary layer of fluid, where the
temperature changes rapidly from its value in the sound wave to the wall
temperature. The presence of large temperature gradients, however, results
in a large dissipation of energy by thermal conduction. For a similar reason,
the fluid viscosity leads to strong absorption of sound when the wave is
incident in an oblique direction. In this case the fluid velocity in the wave
(in the direction of propagation) has a nonzero component tangential to the
surface. At the surface itself, however, the fluid must completely "adhere".
Hence a large tangentialvelocity gradientf must occur in the boundary layer
of fluid, resulting in a large viscous dissipation of energy (see Problem 1).
PROBLEMS
Problem 1. Determine the fraction of energy that is absorbed when a sound wave is
reflected from a solid wall. The density of the wall is supposed so large that the sound does
not penetrate it, and the specific heat so large that the temperature of the wall may be supposed
constant.
Solution. We take the plane of the wall as the plane x = 0, and the plane of incidence as
the xyplane. Let the angle of incidence (which equals the angle of reflection) be 0. The
change in density in the incident wave at any given point on the surface (x = y = 0, say)
is p\ = Ae*~ i(ot . The reflected wave has the same amplitude, so that p\ = p\ at the wall.
The actual change in the fluid density, since both waves (incident and reflected) are propaga
ted simultaneously, is p' = 2Ae~ tu>t . The fluid velocity in the wave is given by v x = cp'iajp,
v 2 = cp'jckjp. The total velocity on the wall, v = Vi+v 2 , is therefore v = v v = 2 A sin x
ceiut/p (or, more precisely, this is what the velocity is found to be when the correct boundary
conditions at the wall in the presence of viscosity are not applied). The actual variation of the
velocity v y along the wall is determined by formula (24.13), and the energy dissipation due to
viscosity by formula (24.14), in which the above expression for v must be substituted for
v e~ ia,t .
The deviation T" of the temperature from its mean value (which is the temperature of the
wall), if calculated without using the correct boundary conditions at the wall, would be found
to be (see (63.13)) T — 2Ac i T^e ib)t tc v p. In reality, however, the temperature distribution is
determined by the equation of thermal conduction, with the boundary condition T" = for
# = 0, and is accordingly given by a formula entirely similar to (24.13).
On calculating the energy dissipation due to thermal conduction as the first term in formula
(77.1), we obtain for the total energy dissipation per unit area of the wall
•femech —
[Vx(^l) + Vvsin^].
P
The mean energy flux density incident on unit area of the wall from the incident wave is
cpVi 2 cos =*= (c a A 2 /2p) cos 0. Hence the fraction of energy absorbed on reflection is
2V(2co)
jvVsin204Vx(— lYI.
c cos 6 L \ c v
This expression is valid only if its value is small (since in deriving it we have supposed the
amplitudes of the incident and reflected waves to be the same). This condition means that
the angle of incidence must not be too near Jir.J
t The normal velocity component is zero at the boundary because of the boundary conditions,
whether or not viscosity is present.
J A calculation of the absorption of sound on reflection at any angle is given by B. P. Konstantinov,
Zhurnal tekhnicheskol fiziki 9, 226, 1939.
§77 Absorption of sound 303
Problem 2. Determine the coefficient of absorption of sound propagated in a cylindrical
pipe.
Solution. The main contribution to the absorption is due to the presence of the walls.
The absorption coefficient y is equal to the energy dissipated at the walls per unit time and
per unit length of the pipe, divided by twice the total energy flux through a crosssection of the
pipe. A calculation entirely similar to that given in Problem 1 leads to the result
[ v , + vx(;i)].
's/co
7= y/2Rc
where R is the radius of the pipe.
Problem 3. Find the dispersion relation for sound propagated in a medium of very high
thermal conductivity.
Solution. In the presence of a large thermal conductivity the flow in a sound wave is not
adiabatic. Hence, instead of the condition of constant entropy, we now have
s=kAT'I p T, (1)
which is the linearised form of equation (49.4) without the viscosity terms. As a second equa
tion we take
P = AP', (2)
which is obtained by eliminating v from equations (63.2) and (63.3). Taking as the funda
mental variables p' and T", we write p' and s' in the form
p' = (dpldT) p T'+(8pldp) T p', s' = (dsldT) p T + (8sldp) T p\
We substitute these expressions in (1) and (2), and then seek T and p' in a form proportional
to e i{ ~ kx ~ m) . The compatibility condition for the resulting two equations for p' and T can
(by using various relations between the derivatives of thermodynamic quantities) be brought
to the form
(or ico\
ct 2 X i
toy
*4_*2 +_+—=<>, (3)
XCs 2
which gives the required relation between k and <a. We have here used the notation
c s 2 = (8pldp)s, ct 2 = (dpl8p) T = c s 2 /y,
where y = cjc v is the ratio of specific heats.
In the limiting case of small frequencies (w <^ c 2 /x)» equation (3) gives
CO C0 2 y / 1 1 \
k = — + i—^l ,
C s 2c s \Ct 2 C s 2 J
which corresponds to the propagation of sound with the ordinary "adiabatic" velocity c t
and a small absorption coefficient which is the second term in (77.6). This is as it should be,
since the condition to <^ c 2 /x means that, during one period, heat can be transmitted only over
a distance r>J V(xI Wl ) (cf. (51.7)) which is small compared with the wavelength c/co.
In the opposite limiting case of large frequencies, we find from (3)
co ct
k = — + i — (c s 2 c T 2 ).
c T 2xc s 2
In this case the sound is propagated with the "isothermal" velocity ct, which is always less
304 Sound §78
than c s . The absorption coefficient is again small compared with the reciprocal of the wave
length, and is independent of the frequency and inversely proportional to the thermal con
ductivity, f
Problem 4. Determine the additional absorption, due to diffusion, of sound propagated
in a mixture of two substances (I. G. Shaposhnikov and Z. A. Gol'dberg 1952).
Solution. The mixture contains an additional source of absorption of sound because the
temperature and pressure gradients occurring in the sound wave result in irreversible pro
cesses of thermal diffusion and barodiffusion (but there is evidently no massconcentration
gradient, and therefore no mass transfer). This absorption is given by the term
(llT P D)(dn/8C) PtT j i*dV
in the rate of change of entropy (58.13) ; we here denote the concentration by Cto distinguish
it from c, the velocity of sound. The diffusion flux is
i =  P D[(k T /T) grad T+ (k p /p) gradp],
with k„ given by (58.10). A calculation similar to that given in §77, using various relations
between the derivatives of thermodynamic quantities, leads to the result that there must be
added to the expression (77.6) for the absorption coefficient a term
Da 2 it dp \ k T l dp \ l dfi \ ) 2
y D _ II I I II II
2c p 2 (8ix/8C) Pi
Problem 5. Determine the effective crosssection for the absorption of sound by a sphere
of radius small compared with • v / ('V a, )•
Solution. The total absorption is composed of the effects of the viscosity and thermal
conductivity of the gas. The former is given by the work done by the Stokes frictional force
when gas moving in a sound wave flows round a sphere ; as in §76, Problem 3, it is assumed that
the sphere is not moved by this force. The effect of conductivity is given by the amount of
heat q transferred from the gas to the sphere per unit time ( §76, Problem 3) : the energy dissi
pation when an amount of heat q is transferred, the temperature difference between the gas
(far from the sphere) and the sphere being T, is qT'jT. The total effective absorption cross
section is found to be
2ttR
c
H^ 1 )]
§78. Second viscosity
The second viscosity coefficient £ (which we shall call simply the second
viscosity) is usually of the same order of magnitude as the viscosity coefficient
7], There are, however, cases where £ can take values considerably exceeding
rj. As we know, the second viscosity appears in processes which are accom
panied by a change in volume (i.e. in density) of the fluid. In compression
or expansion, as in any rapid change of state, the fluid ceases to be in thermo
dynamic equilibrium, and internal processes are set up in it which tend to
f The second root of equation (3), which is quadratic in k 2 , corresponds to "thermal waves" which
are rapidly damped with increasing x. In the limit a>x <^ c 2 this root gives
* = V(Wx) = (i+0V(W2x),
in agreement with (52.17). In the case cax ^> c 2 we have
k = (1 +i)\/(cocv/2xc p ).
§78 Second viscosity 305
restore this equilibrium. These processes are usually so rapid (i.e. their relaxa
tion time is so short) that the restoration of equilibrium follows the change
in volume almost immediately unless, of course, the rate of change of volume
is very large.
It may happen, nevertheless, that the relaxation times of the processes of
restoration of equilibrium are long, i.e. they take place comparatively slowly.
For instance, if we are concerned with a liquid or gas which is a mixture of
substances between which a chemical reaction occurs, there is a state of chemi
cal equilibrium, characterised by the concentrations of the substances in
the mixture, for any given density and temperature. If, for example, we
compress the fluid, the state of equilibrium is destroyed, and a reaction
begins, as a result of which the concentrations of the substances tend to take
the equilibrium values corresponding to the new density and temperature.
If this reaction is not rapid, the restoration of equilibrium occurs relatively
slowly and does not immediately follow the compression. The latter process
is then accompanied by internal processes which tend towards the equilibrium
state. But the processes which establish equilibrium are irreversible; they
increase the entropy, and therefore involve energy dissipation. Hence, if the
relaxation time of these processes is long, a considerable dissipation of energy
occurs when the fluid is compressed or expanded, and, since this dissipation
must be determined by the second viscosity, we reach the conclusion that
£ is large.f
The intensity of the dissipative processes, and therefore the value of £,
depend of course on the relation between the rate of compression or expansion
and the relaxation time. If, for example, we have compression or expansion
due to a sound wave, the second viscosity will depend on the frequency of the
wave. Thus the second viscosity is not just a constant characteristic of the
material concerned, but depends on the frequency of the motion in which it
appears. The dependence of £ on the frequency is called its dispersion.
The following general method of discussing all these phenomena is due to
L. I. Mandel'shtam and M. A. Leontovich (1937). Let £ be some physical
quantity characterising the state of a body, and o its value in the equilibrium
state; o is a function of density and temperature. For instance, in fluid mix
tures  may be the concentration of one component, and then o is the con
centration in chemical equilibrium.
If the body is not in equilibrium, £ will vary with time, tending to the value
o In states close to equilibrium the difference £ — £o is small, and we can
expand the rate of change £ of £ in a series of powers of this difference.
The zeroorder term is absent, since £ must be zero in the equilibrium state,
i.e. when £ = £o Hence, as far as the firstorder term, we have
£= (£&)/t. (78.1)
The proportionality coefficient must be negative, since otherwise £ would not
f A slow process which results in a large £ is often the transfer of energy from translator/ degrees
of freedom of a molecule to vibrational (intramolecular) degrees of freedom.
306 Sound §78
tend to a finite limit. The positive constant t is of the dimensions of time,
and may be regarded as the relaxation time for the process in question; the
greater is t, the more slowly the approach to equilibrium takes place.
In what follows we shall consider processes in which the fluid is subjected
to a periodic adiabaticf compression and expansion, so that the variable part
of the density (and of the other thermodynamic quantities) depends on the
time through a factor e i<ot ; we are considering a sound wave in the fluid.
Together with the density and other quantities, the position of equilibrium
also varies, so that £ can be written as £ = £oo+fo', where £00 is the
constant value of £ corresponding to the mean density, and £o' is a periodic
part, proportional to e~ iut . Writing the true value £ in the form g = £ o+ I',
we conclude from equation (78.1) that £' also is a periodic function of time,
related to £ o' by
£' = So'Klian). (78.2)
Let us calculate the derivative of the pressure with respect to the density
for the ptocess in question. The pressure must now be regarded as a function
of the density and of the value of £ in the state concerned, and also of the
entropy, which we suppose constant and, for brevity, omit. Then
dpi dp = (dpldp\+(dpldt) p 8£/d P .
In accordance with (78.2), we substitute here
% %' 1 a&' 1 a&
obtaining
dp dp I — tear dp 1 — icoT dp
»._L(*) + (»)*.U»)].
dp lton\\dp/ i \d£/ p dp \8pf s l
The sum (dpl8p)g + (dpld$) p d(joldp is just the derivative of p with respect to
p for a process which is so slow that the fluid remains in equilibrium; denoting
it by (dp}dp) m , we have finally
(78.3)
dp 1 — iorr _ \ dp} eq \ dp J g_
Next, let po be the pressure in a state of thermodynamic equilibrium;
Po is related to the other thermodynamic quantities by the equation of state
of the fluid, and is entirely determined when the density and entropy are
given. The pressure p in a nonequilibrium state, however, differs from po,
and is a function of £ also. If the density is adiabatically increased by 8p,
the equilibrium pressure changes by Spo = {dpjdp) eq hp i while the total
increase in the pressure is (dp/8p)Sp, with dp/dp given by formula (78.3).
t The change in the entropy (in states close to equilibrium) is of the second order of smallness.
Hence, to this order of accuracy, we can speak of an adiabatic process.
§78 Second viscosity 307
Hence the difference p — />o between the true pressure and the equilibrium
pressure, in a state where the density is p + 8p, is
I dp \dp /eqj 1 — icor \_\ dp / eq \ dp / ^J
We are here interested in the density changes due to the motion of the
fluid. Then 8p is related to the velocity by the equation of continuity,
which we write in the form d(8p)jdt+p div v = 0, where d/dt denotes the
total time derivative. In a periodic motion we have d(8p)/dt = — ico8p,
and therefore 8p = (pjico) div v. Substituting this expression in (78.3a),
we obtain
PPo = t^t (co 2 cJ) divv, (78.4)
1— 10)T
where we have used the notation
co 2 = (dpldp)^ cj = (dpld P ) g , (78.5)
the significance of which will be explained below.
In order to relate these expressions to the viscosity of the fluid, we write
down the stress tensor aye. In this tensor the pressure appears in the term
— p8iic. Subtracting the pressure po determined by the equation of state, we
find that in a nonequilibrium state o^ contains an additional term
rp
{ppo)8ik = — L . — (cjco^ucdivv.
\—l(DT
Comparing this with the general expression (15.2) and (15.3) for the stress
tensor, in which div v appears in the term £ div v, we conclude that the
presence of slow processes tending to establish equilibrium is macroscopically
equivalent to the presence of a second viscosity given by
t = Tp(cJc *)l(licoT). (78.6)
These processes do not affect the ordinary viscosity 7). For processes so slow
that cot <^ 1, £ is
£o = t P (cJc 2); (78.7)
it increases with the relaxation time r, in accordance with what was said
above. For large frequencies, £ depends on the frequency, i.e. it exhibits
dispersion.
Let us now consider the question of how the presence of processes with
large relaxation times (for definiteness, we shall speak of chemical reactions)
affects the propagation of sound in a fluid. To do so, we might start from
the equation of motion of a viscous fluid, with £ given by formula (78.6).
It is simpler, however, to consider a motion in which viscosity is neglected
but the pressure p is given by the above formulae instead of by the equation
of state. The general relations which we obtained in §63 then remain formally
applicable. In particular, the wave number and the frequency are still
308 Sound §78
related by k = co/c, where c = ^/(dpjdp), and the derivative dp/dp is now
given by (78.3) ; the quantity c, however, no longer denotes the velocity of
sound, being complex. Thus we obtain
k = coV[(1^)/(c 2 Coo 2 *wt)]. (78.8)
The "wave number" given by this formula is complex. The meaning of
this fact is easily seen. In a plane wave, all quantities depend on the co
ordinate x (the #axis being in the direction of propagation) through a factor
e ikx . Writing k in the form k = ki + ik 2 with k\, k 2 real, we have e ikx =
e i\x e Jc 2 x t i #e besides the periodic factor e ik i x we have a damping factor e~ k 2 x
(&2 mustj of course, be positive). Thus the complex nature of the wave
number formally expresses the fact that the wave is damped, i.e. there is
absorption of sound. The real part of the complex wave number gives the
variation in phase of the wave with distance, and the imaginary part is the
absorption coefficient.
It is not difficult to separate the real and imaginary parts of (78.8). In
the general case of arbitrary co the expressions for k\ and k 2 are rather cum
bersome, and we shall not write them out here. It is important that k\
is a function of the frequency (as is £2) Thus, if chemical reactions can occur
in the fluid, the propagation of sound at sufficiently high frequencies is
accompanied by dispersion.
In the limiting case of low frequencies (cot <^ 1), formula (78.8) gives
to a first approximation k = cojco, corresponding to the propagation of sound
with velocity cq. This is as it should be, of course: the condition cot <4 1
means that the period 1/co of the sound wave is large compared with the
relaxation time, i.e. the establishment of chemical equilibrium follows the
variations of density in the sound wave, and the velocity of sound is deter
mined by the equilibrium value of the derivative dp/ dp. In the second approxi
mation We have
co ico 2 T
k =  + —(cJco*), (78.9)
CO lC{f
i.e. damping occurs, with a coefficient proportional to the square of the fre
quency. Using (78.7), we can write the imaginary part of k in the form
k% = co 2 £o/2pco 3 ; this agrees with the ^dependent part of the absorption
coefficient y as given by (77.6), which was obtained without taking account
of the dispersion.
In the opposite limiting case of high frequencies (cot > 1), we have
in the first approximation k = cojc^, i.e. the propagation of sound with
velocity c ro — again a natural result, since for cot p 1 we can suppose that
no reaction occurs during a single period, and the velocity of sound must
therefore be determined by the derivative (dpjdp)g taken at constant concen
tration. The second approximation gives
k =  + i^—±. (78.10)
c^ 2tc*
§78 Second viscosity 309
The damping coefficient is independent of the frequency. As we go from
co <^ 1/r to o) ^> 1/t, this coefficient increases monotonically to the constant
value given by formula (78.10). It should be noted that the quantity A 2 /&i,
which represents the amount of absorption over a distance of one wavelength,
is small in both limiting cases {fa\k\ < 1) ; it has a maximum at some inter
mediate frequency, namely co = v/OVO/ 7 "
It is seen from (78.7) (e.g.) that
Coo > co, (78.11)
since we must have £ > 0. The same result can be obtained by simple
arguments based on Le Chatelier's principle. Let us suppose that the
volume of the system is reduced, and the density increased, by some external
agency. The system is thereby brought out of equilibrium, and according
to Le Chatelier's principle processes must begin which tend to reduce the
pressure. This means that dp/ dp will decrease, and, when the system returns
to equilibrium, the value of dpi dp = c 2 will be less than in the nonequili
brium state.
In deriving all the above formulae we have assumed that there is only a
single slow internal process of relaxation. Cases are also possible where
several different such processes occur simultaneously. All the formulae can
easily be generalised to cover such cases. Instead of a single quantity £,
we now have several quantities £i, £2, ... which characterise the state of the
system, and a corresponding series of relaxation times ti, T2, .... We choose
the quantities £ w in such a way that each of the derivatives  n depends
only on the corresponding », i.e. so that
in = tf»f«o)/T„. (78.12)
Calculations entirely similar to the above then give
& = cJ+J^ a n l(licoT n ), (78.13)
n
where c ro 2 = {dpjdp)^ and the constants a n are
a n = (Bpld€nW£nldp)e+ ( 78  14 )
If there is only one quantity £, formula (78.13) becomes (78.3), as it should.
CHAPTER IX
SHOCK WAVES
§79. Propagation of disturbances in a moving gas
When the velocity of a fluid in motion becomes comparable with or exceeds
that of sound, effects due to the compressibility of the fluid become of prime
importance. Such motions are in practice met with in gases. The dynamics
of highspeed flow is therefore usually called gas dynamics.
It should be mentioned first of all that, in gas dynamics, the Reynolds
numbers involved are almost always very large. For the kinematic viscosity
of a gas is, as we know from the kinetic theory of gases, of the order of the
mean free path / of the molecules multiplied by the mean velocity of their
thermal motion; the latter is of the same order as the velocity of sound, so that
v ~ cl. If the characteristic velocity in a problem of gas dynamics is also of
the order of c, then the Reynolds number R ~ Lc\v ~ Ljl, i.e. it is deter
mined by the ratio of the dimension L to the mean free path /, which we know
is very large.f As always occurs when R is very large, the viscosity has an
important effect on the motion of the gas only in a very small region, and in
what follows we shall (except where the contrary is specifically stated) regard
the gas as an ideal fluid.
The flow of a gas is entirely different in nature according as it is subsonic
or supersonic, i.e. the velocity is less than or greater than that of sound.
One of the most important distinctive features of supersonic flow is the fact
that there can occur in it what are called shock waves, whose properties we
shall examine in detail in the following sections. Here we shall consider
another characteristic property of supersonic flow, relating to the manner of
propagation of small disturbances in the gas.
If a gas in steady motion receives a slight perturbation at any point, the
effect of the perturbation is subsequently propagated through the gas with
the velocity of sound (relative to the gas itself). The rate of propagation of
the disturbance relative to a fixed system of coordinates is composed of
two parts: firstly, the perturbation is "carried along" by the gas flow with
velocity v and, secondly, it is propagated relative to the gas with velocity c
in any direction n. Let us consider, for simplicity, a uniform flow of
gas with constant velocity v, subjected to a small perturbation at some
point O (fixed in space). The velocity v+cn with which the perturbation
is propagated from O (relative to the fixed system of coordinates) has
different values for different directions of the unit vector n. We obtain
f We shall not consider the problem of the motion of bodies in very rarefied gases, where the
mean free path of the molecules is comparable with the dimension of the body. This problem is in
essence not one of fluid dynamics, and must be examined by means of the kinetic theory of gases.
310
§79 Propagation of disturbances in a moving gas 311
all its possible values by placing one end of the vector v at the point O
and drawing a sphere of radius c centred at the other end. The vectors from
O to points on this sphere give the possible magnitudes and directions of the
velocity of propagation of the perturbation. Let us first suppose that v < c .
Then the vector v + cn can have any direction in space (Fig. 40a). That
is, a disturbance which starts from any point in a subsonic flow will eventually
reach every point in the gas. If, on the other hand, v > c, the direction of the
vector v + cn can lie, as we see from Fig. 40b, only in a cone with its vertex at
O, which touches the sphere with its centre at the other end of the vector v.
If the aperture of the cone is 2a, then, as is seen from the figure,
sin a = cjv. (79.1)
Fig. 40
Thus a disturbance starting from any point in a supersonic flow is propagated
only downstream within a cone whose aperture is the smaller, the smaller the
ratio cjv. A disturbance starting from O does not affect the flow outside
this cone.
The angle a determined by equation (79.1) is called the Mach angle.
The ratio v/c itself, which often occurs in gas dynamics, is the Mach number M :
M = v/c. (79.2)
The surface bounding the region reached by a disturbance starting from a
given point is called the Mach surface or characteristic surface.
In the general case of an arbitrary steady flow, the Mach surface is not a
cone throughout the volume. However, it can be asserted that, as before, this
surface cuts the streamline through any point on it at the Mach angle. The
value of the Mach angle varies from point to point with the velocities v and c.
It should be emphasised here, incidentally, that, in flow with high velocities,
the velocity of sound is different at different points : it varies with the ther
modynamic quantities (pressure, density, etc.) of which it is a function. f
The velocity of sound as a function of the coordinates is sometimes called
the local velocity of sound.
t In the discussion of sound waves given in Chapter VIII, the velocity of sound could be regarded
as constant.
312 Shock Waves §79
The properties of supersonic flow described above give it a character quite
different from that of subsonic flow. If a subsonic gas flow meets any
obstacle (if, for instance, it flows past a body), the presence of this obstacle
affects the flow in all space, both upstream and downstream; the effect
of the obstacle is zero only asymptotically at an infinite distance from it.
A supersonic flow, however, is incident "blindly" on an obstacle; the effect
of the latter extends only downstream,f and in all the remaining part of
space upstream the gas flows as if the obstacle were absent.
In the case of steady plane flow of a gas, the characteristic surfaces can be
replaced by characteristic lines (or simply characteristics) in the plane of the
flow. Through any point O in this plane there pass two characteristics (AA'
and BB' in Fig. 41), which intersect the streamline through this point at the
Mach angle. The downstream branches OA and OB of the characteristics
maybe said to leave the point O; they bound the region AOB of the flow
where perturbations starting from O can take effect. The branches B'O
and A'O may be said to reach the point O; the region A' OB' between them
is that which can affect the flow at O.
The concept of characteristics (surfaces in the threedimensional case)
has also a somewhat different aspect. They are rays along which disturbances
are "propagated" which satisfy the conditions of geometrical acoustics. If,
for example, a steady supersonic gas flow meets a fairly small obstacle, then a
steady perturbation of the gas flow will be found along the characteristics
which leave this obstacle. The same result was reached in §67 from a study
of the geometrical acoustics of moving media.
When we speak of a perturbation of the state of the gas, we mean a slight
change in any of the quantities characterising its state : the velocity, pressure,
t To avoid misunderstanding, we should mention that, if a shock wave is formed in front of the
obstacle, this region is somewhat enlarged (see §114).
§80 Steady flow of a gas 313
density, etc. The following remark should be made on this point. Pertur
bations in the values of the entropy of the gas (for constant pressure) and of
its vorticity are not propagated with the velocity of sound. These perturba
tions, once having arisen, do not move relative to the gas; relative to a fixed
system of coordinates they move with the gas at the velocity appropriate to
each point. For the entropy, this is an immediate consequence of the law of
conservation (in an ideal fluid),
ds/dt = ds/dt+vgrads = 0,
which shows that the entropy of any given volume element in the gas remains
constant as the element moves about, i.e. each value of s moves with the
point to which it belongs. The same result for the vorticity follows from the
conservation of circulation.
Thus we can say that, for perturbations of entropy and vorticity, the
characteristics are the streamlines. This, of course, does not affect the general
validity of the statements made above about regions of influence, since
they were based only on the existence of a maximum velocity of propagation
(that of sound) of disturbances relative to the gas itself.
§80. Steady flow of a gas
We can obtain immediately from Bernoulli's equation a number of general
results concerning adiabatic steady flow of a gas. The equation is, for steady
flow, w + \v 2 = constant along each streamline; if we have potential flow,
then the constant is the same for every streamline, i.e. at every point in the
fluid. If there is a point on some streamline at which the gas velocity is zero,
then we can write Bernoulli's equation as
w+±v 2 = wo, (80.1)
where wq is the value of the heat function at the point where v = 0.
The equation of conservation of entropy for steady flow is v«grads
= vdsjdl = 0, i.e. 5 is constant along each streamline. We can write this in a
form analogous to (80.1):
s = s . (80.2)
We see from equation (80.1) that the velocity v is greater at points where
the heat function w is smaller. The maximum value of the velocity (on
the streamline considered) is found at the point where w is least. For con
stant entropy, however, we have dw = dpjp\ since p > 0, the differentials
dw and dp have like signs, and therefore w and p vary in the same sense.
We can therefore say that the velocity increases along a streamline when the
pressure decreases, and vice versa.
The smallest possible values of the pressure and the heat function (in
adiabatic flow) are obtained when the absolute temperature T = 0. The
corresponding pressure is p = 0, and the value of w for T = can be
arbitrarily taken as the zero of energy; then w = for T = 0. We can
314
Shock Waves
§80
now deduce from (80.1) that the greatest possible value of the velocity (for
given values of the thermodynamic quantities at the point where v = 0) is
%ax = V(2wo). (80.3)
This velocity can be attained when a gas flows steadily out into a vacuum.f
Let us now consider how the mass flux density j = pv varies along a
streamline. From Euler's equation (vgrad)v = (l//>)grad/>, we find
that the relation v dv = dpjp between the differentials dv and dp holds
along a streamline. Putting dp = c 2 dp, we have
dp/dv = —pvjc 2
and, substituting in d(pv) = p dv + v dp, we obtain
d(pv){dv = p(l—v 2 /c 2 ).
(80.4)
(80.5)
tf.
079
O50
025
025 050 075
TOO
125
Y/
Fig. 42
150
175 200 225 250
From this we see that, as the velocity increases along a streamline, the
mass flux density increases as long as the flow remains subsonic. In the super
sonic range, however, the mass flux density diminishes with increasing
velocity, and vanishes together with p when v — v ma , x (Fig 42). This im
portant difference between subsonic and supersonic steady flows can be simply
interpreted as follows. In a subsonic flow, the streamlines approach in the
direction of increasing velocity. In a supersonic flow, however, they diverge
in that direction.
The flux j has its maximum value j+ at the point where the gas velocity is
equal to the local velocity of sound :
j* = p*c*, (80.6)
where the asterisk suffix indicates values corresponding to this point. The
t In reality, of course, when there is a sharp fall in temperature the gas must condense and form a
twophase "fog". This, however, does not essentially affect the results given.
§80 Steady flow of a gas 315
velocity v* = c # is called the critical velocity. In the general case of an arbi
trary gas, the critical values of quantities can be expressed in terms of their
values at the point v = 0, by solving the simultaneous equations
s* = % «> # + c* 2 = w . (80.7)
It is evident that, whenever M = vjc < 1, we have also vfc* < 1, and
if M > 1 then vjc^ > 1. Hence the ratio M* = v\c % serves in this case as a
criterion analogous to M, and is more convenient, since c % is a constant,
unlike c, which varies along the stream.
In applications of the general equations of gas dynamics, the case of a
perfect gas is of particular importance. For a perfect gas we know from
thermodynamics all the relations between the various thermodynamic
quantities, and these relations are very simple. This makes it possible to give
a complete solution of the equations of gas dynamics in many cases.
We shall give here, for reference, the relations between the various thermo
dynamic quantities for a perfect gas, since they will often be needed in what
follows. We shall always assume (unless otherwise stated) that the specific
heat of a perfect gas is independent of temperature.
The equation of state for a perfect gas is
pV = pip = RT/p, (80.8)
where R = 8314 xlO 7 erg/deg is the gas constant, and /x the molecular
weight of the gas. The velocity of sound in a perfect gas is, as shown in §63,
given by
c* = yRT/fi = yp/p, (80.9)
where we have introduced the constant ratio of specific heats y = c v Jc v ,
which always exceeds unity; for monatomic gases y = 5/3, and for diatomic
gases y — 7/5, at ordinary temperatures.
The internal energy of a perfect gas is, apart from an unimportant additive
constant,
c = c v T = pV\iy 1) = c*ly{y 1). (80.10)
For the heat function we have the analogous formulae
to = CpT = ypV/(y 1) = c^{y 1). (80.11)
Here we have used the wellknown relation c p — c v = R/p,. Finally, the
entropy of the gas is
s = c v log(pl P r) = c v log(pVr/p). (80.12)
Let us now investigate steady flow, applying the general relations pre
viously obtained to the case of a perfect gas. Substituting (80.11) in (80.3),
we find that the maximum velocity of steady flow is
%ax = coV[2/(yl)l (80.13)
316 Shock Waves §80
For the critical velocity we obtain from the second equation (80.7)
+ \c£ = Wq =
y\ y\
whencej*
c* = W[2/(y+l)]. (80.14)
Bernoulli's equation (80.1), after substitution of the expression (80.11)
for the heat function, gives the relation between the temperature and the
velocity at any point on the streamline ; similar relations for the pressure and
density can then be obtained directly by means of Poisson's adiabatic equa
tion:
(80.15)
Thus
p = P o(777o)i/Cri), p =
we obtain the important results
rr.[i« y i)Jr.(ir
r v 2 i l/(rD /
p = po[iKri)^J =po(i
/>=/»o[iKri)^J =Po(i
Po(p/po) y .
1 V 2 \
•
+ i c.*r
y—\ v 2
y+i a
y— 1 v 2
y+\ c 2
\l/(yl)
■) •
.)■•]
(80.16)
It is sometimes convenient to use these relations in a form which gives the
velocity in terms of other quantities :
^ = iLALf, _(£pn . ^thiJLfX (8 o.i7)
y1 po L \po/ J y—lpol \ po 1 J
We may also give the relation between the velocity of sound and the
velocity v :
C 2 = co 2 ±(y \)v 2 = ±(y + l)*. a i(y l)^ 2 . (80.18)
Hence we find that the numbers M and M* are related by
y+1
M*2 = L . ; (80.19)
yl+2/M2' V ;
when M varies from to oo, M* 2 varies from to (y+ l)/(y— 1).
Finally^ we may give expressions for the critical temperature, pressure and
density: they are obtained by putting v = c% in formulae (80.16)$:
f Fig. 42 shows the ratio jlj# as a function of vjc^ for air (y = 14, f ma x = 245^).
J For air, e.g., (y = 14)
c* = 0913co, p* = 0528/>o, />* = 0634/j , T m = 0833 T .
§81 Surfaces of discontinuity 317
T* = 27o/(y+l),
/ 2 \r/(ri)
/>* = po — — 
\y+l
(80.20)
In conclusion, it should be emphasised that the results derived above are
valid only for flow in which shock waves do not occur. When shock waves are
present, equation (80.2) does not hold ; the entropy of the gas increases when
a streamline passes through a shock wave. We shall see, however, that
Bernoulli's equation (80.1) remains valid even when there are shock waves,
since w + %v 2 is a quantity which is conserved across a surface of discon
tinuity (§82); formula (80.14), for example, therefore remains valid also.
PROBLEM
Express the temperature, pressure and density along a streamline in terms of the Mach
number.
Solution. Using the formulae obtained above, we find
T /T = 1 +Ky 1)M 2 , po/p = [1 +1(7 l)M2]7/(7i),
poIp = [l+Mr^M 2 ] 1 ^ 1 ).
§81. Surfaces of discontinuity
In the preceding chapters we have considered only flows such that all
quantities (velocity, pressure, density, etc.) vary continuously. Flows are
also possible, however, for which discontinuities in the distribution of these
quantities occur.
A discontinuity in a gas flow occurs over one or more surfaces ; the quan
tities concerned change discontinuously as we cross such a surface, which is
called a surface of discontinuity. In nonsteady gas flow the surfaces of dis
continuity do not in general remain fixed; here it should be emphasised,
however, that the rate of motion of these surfaces bears no relation to the
velocity of the gas flow itself. The gas particles in their motion may cross a
surface of discontinuity.
Certain boundary conditions must be satisfied on surfaces of discontinuity.
To formulate these conditions, we consider an element of the surface and use
a coordinate system fixed to this element, with the #axis along the normal.f
Firstly, the mass flux must be continuous : the mass of gas coming from
one side must equal the mass leaving the other side. The mass flux through
the surface element considered is pv x per unit area. Hence we must have
pivix = P2V2x, where the suffixes 1 and 2 refer to the two sides of the surface
of discontinuity.
f If the flow is not steady, we consider an element of the surface during a short interval of time.
318 Shock Waves §81
The difference between the values of any quantity on the two sides of
the surface will be denoted by enclosing it in square brackets; for example,
[pv x ] == piviz— p2V2z, and the condition just derived can be written
[f>vx] = 0. (81.1)
Next, the energy flux must be continuous. The energy flux is given by
(6.3). We therefore obtain the condition
[pvafttP + w)] = 0. (81.2)
Finally, the momentum flux must be continuous, i.e. the forces exerted
on each other by the gases on the two sides of the surface of discontinuity
must be equal. The momentum flux per unit area is (see §7) pm + pViVtfik.
The normal vector n is along the rcaxis. The continuity of the ^component
of the momentum flux therefore gives the condition
Ip+pvJ] = 0, (81.3)
while that of the y and z components gives
IpVaPy] = 0, \pv x v z ] = 0. (81.4)
Equations (81.1)— (81.4) form a complete system of boundary conditions
at a surface of discontinuity. From them we can immediately deduce the
possibility of two types of surface of discontinuity.
In the first type, there is no mass flux through the surface. This means
that pivix = p$V2x = 0. Since pi and p2 are not zero, it follows that v\ x
= V2x = 0. The conditions (81.2) and (81.4) are then satisfied, and the con
dition (81.3) gives pi = P2. Thus the normal velocity component and the
gas pressure are continuous at the surface of discontinuity:
vix = v 2x = 0, \p] = 0, (81.5)
while the tangential velocities v y , v z and the density (as well as the other
thermodynamic quantities except the pressure) may be discontinuous by any
amount. We call this a tangential discontinuity.
In the second type, the mass flux is not zero, and v\ x and V2x are therefore
also not zero. We then have from (81.1) and (81.4)
[vy] = 0, [v z ] = 0, (81.6)
i.e. the tangential velocity is continuous at the surface of discontinuity.
The pressure, the density (and the other thermodynamic quantities) and the
normal velocity, however, are discontinuous, their discontinuities being
related by (81.1) — (81.3). In the condition (81.2) we can cancel pv x by
(81.1), and replace v 2 by v x 2 since v y and v z are continuous. Thus the
following conditions must hold at the surface of discontinuity in this case :
b*>x] = 0, \
[W + «7] = 0, (81.7)
[p + pv x *] = 0. J
A discontinuity of this kind is called a shock wave, or simply a shock.
§82 The shock adiabatic 319
If we now return to the fixed coordinate system, we must everywhere
replace v x by the difference between the gas velocity component v n normal
to the surface of discontinuity and the velocity u of the surface itself, which is
defined to be normal to the surface :
v x = v n — u. (81.8)
The velocities v and « are taken in the fixed system. The velocity v% is
the velocity of the gas relative to the surface of discontinuity; we can also
say that — v x = u — v n is the rate of propagation of the surface relative to
the gas. It should be noticed that, if v x is discontinuous, this velocity has
different values relative to the gas on the two sides of the surface.
We have already discussed (in §30) tangential discontinuities, at which the
tangential velocity component is discontinuous, and we showed that, in an
incompressible fluid, such discontinuities are absolutely unstable and must
result in a turbulent region. A similar investigation for a compressible fluid
shows that the same instability occurs, for any velocities.
A particular "degenerate" case of tangential discontinuity is that where the
velocity is continuous, but not the density (and therefore the other thermo
dynamic quantities, except the pressure). The above remarks on instability
do not relate to discontinuities of this kind.
§82. The shock adiabatic
Let us now investigate shock waves in detail. We have seen that, in this
type of discontinuity, the tangential component of the gas velocity is con
tinuous. We can therefore take a coordinate system in which the surface
element considered is at rest, and the tangential component of the gas velocity
is zero on both sides.f Then we can write the normal component v x as v
simply, and the conditions (81.7) take the form
pivi = p 2 v 2 =j, (82.1)
pi + pivi 2 = pz + P2^2 2 , (82.2)
Wl + ©l 2 = W2 + ^2 2 , (82.3)
where j denotes the mass flux density at the surface of discontinuity. In
what follows we shall always take j positive, with the gas going from side 1
to side 2. That is, we call gas 1 the one into which the shock wave moves,
and gas 2 that which remains behind the shock. We call the side of the shock
wave towards gas 1 the front of the shock, and that towards gas 2 the back.
We shall derive a series of relations which follow from the above condi
tions. Using the specific volumes V\ = l//>i, V2 = I//02, we obtain from
(82.1)
vi = jV lt v 2 = jV 2 (82.4)
t This coordinate system is used everywhere in §§8285, 87, 88.
A shock wave at rest is called a compression discontinuity. If the shock is perpendicular to the
direction of flow, we have a normal shock, otherwise an oblique shock.
320 Shock Waves
and, substituting in (82.2),
pi+pVi = p2+j 2 V 2 ,
§82
(82.5)
or
j 2 = (p2pi)l(V 1 V 2 ). (82.6)
This formula, together with (82.4), relates the rate of propagation of a shock
wave to the pressures and densities of the gas on the two sides of the surface.
Fig. 43
Since p is positive, we see that either p% > pi, V\ > V2, or p% < pi,
V\ < V2', we shall see below that only the former case can actually occur.
We may note the following useful formula for the velocity difference
©1 — ^2. Substituting (82.6) in V1 — V2 = j{V\— V2), we obtainf
Vl V2 = V[(p2piWiV2)]. (82.7)
Next, we Write (82.3) in the form
^l+i/W = W2 + hpV 2 2 (82.8)
and, substituting/ 2 from (82.6), obtain
W!w 2 + !( V x + V 2 ){p 2 pi) = 0. (82.9)
If we replace the heat function why e+pV, where e is the internal energy, we
can write this relation as
eie2 + l(V 1 V 2 ){pi+p2) = 0.
(82.10)
These relations hold between the thermodynamic quantities on the two sides
of the surface of discontinuity.
For given p±, V\, equation (82.9) or (82.10) gives the relation between p2
and V 2 . This relation is called the shock adiabatic or the Hugoniot adiabatic
(W. J. M. Rankine, 1870; H. Hugoniot, 1889). It is represented graphically
in the^Fplane (Fig. 43) by a curve passing through the given point (pi, V{)
f Here we write the positive square root, since, as we shall see later (§84) we must have v x — v 2 > 0.
482
The shock adiabatic
321
(for pi = p2, Vi = V 2 we have also ei = e 2 , so that (82.10) is satisfied identi
cally). It should be noted that the shock adiabatic cannot intersect the vertical
line V = V\ except at (pi, Vi). For the existence of another intersection
would mean that two different pressures satisfying (82.10) correspond to the
same volume. For V\  V 2 , however, we have from (82.10) also ei = e 2 ,
and when the volumes and energies are the same the pressures must be the
same. Thus the line V = V\ divides the shock adiabatic into two parts,
each of which lies entirely on one side of the line. Similarly, the shock
adiabatic meets the horizontal line p = pi only at the point {pi, Vi).
ba
Fig. 44
Let aa' (Fig. 44) be the shock adiabatic through the point (pi, Vi) as a
state of gas 1. We take any point (p2, V 2 ) on it and draw through that point
another adiabatic bb', for which (p 2 , V 2 ) is a state of gas 1. It is evident that
the pair of values (pi, Vi) satisfies the equation of this adiabatic also. The
adiabatics aa' and bb' therefore intersect at the two points (pi, Vi) and (p 2 ,
V 2 ). It must be emphasised that the adiabatics are not identical, as would
happen for Poisson adiabatics through a given point. This is a consequence
of the fact that the equation of the shock adiabatic cannot be written in the
form /(p, V) = constant, where / is some function, whereas the Poisson
adiabatic, for example, can be written s(p, V) — constant. The Poisson
adiabatics for a given gas form a oneparameter family of curves, but the
shock adiabatic is determined by two parameters, the initial values pi and
V\. This has also the following important result: if two (or more) successive
shock waves take a gas from state 1 to state 2 and from there to state 3, the
transition from state 1 to state 3 cannot in general be effected by the passage
of any one shock wave.
For a given initial thermodynamic state of the gas (i.e. for given pi and
Vi), the shock wave is defined by only one parameter; for instance, if the
pressure p 2 behind the shock is given, then Vi is determined by the Hugoniot
adiabatic, and the flux density; and the velocities v\ and v 2 are then given by
formulae (82.4) and (82.6). It should be mentioned, however, that we are
322 Shock Waves §83
here considering the shock wave in a coordinate system in which the gas is
moving normal to the surface. If the shock wave may be situated obliquely
to the direction of flow, another parameter is needed; for example, the value
of the velocity component tangential to the surface.
The following convenient graphical interpretation of formula (82.6)
may be mentioned. If the point (p ly Vi) on the shock adiabatic (Fig. 43)
is joined by a chord to any other point (p 2 , V 2 ) on it, then (p 2 pi)l(V 2 Vi)
= j 2 is Just the slope of this chord relative to the axis of abscissae. Thus;',
and therefore the velocity of the shock wave, are determined at each point
of the shock adiabatic by the slope of the chord joining that point to the
point (pi,; Fi).
Like the other thermodynamic quantities, the entropy is discontinuous at
a shock wave. By the law of increase of entropy, the entropy of a gas can
only increase during its motion. Hence the entropy s 2 of the gas which has
passed through the shock wave must exceed its initial entropy s± :
*2 > si. (82.11)
We shall see below that this condition places very important restrictions on
the manner of variation of all quantities in a shock wave.
The following fact should be emphasised. The presence of shock waves
results in an increase in entropy in those flows which can be regarded as
motions of an ideal fluid in all space, the viscosity and thermal conductivity
being zero. The increase in entropy signifies that the motion is irreversible,
i.e. energy is dissipated. Thus the discontinuities are a means by which
energy can be dissipated in the motion of an ideal fluid. It follows that
d'Alembert's paradox (§11) does not arise when bodies move in an ideal fluid
in such a way as to cause shock waves. In such cases there is a drag force.
The true mechanism by which the entropy increases in shock waves lies,
of course, in dissipative processes occurring in the very thin layers which
actual shock waves are (see §87). It should be noticed, however, that the
amount of this dissipation is entirely determined by the laws of conservation
of mass, energy and momentum, when they are applied to the two sides of
such layers; the width of the layers is just such as to give the increase in en
tropy required by these conservation laws.
The increase in entropy in a shock wave has another important effect on
the motion : even if we have potential flow in front of the shock wave, the flow
behind it is in general rotational. We shall return to this matter in §106.
§83. Weak shock waves
Let us consider a shock wave in which the discontinuity in every quantity
is small; we call this a weak shock wave. We transform the relation (82.9)
by expanding in powers of the small differences *2*i and p 2 pi. We
shall see that the first and secondorder terms in p 2 —pi then cancel; we
must therefore carry the expansion with respect top 2 pi as far as the third
§83 Weak shock waves 323
order. In the expansion with respect to $2  *i, only the firstorder terms need
be retained. We have
W2W1 = (dw/dsi)p(s2si) + (dwl8pi) s (p2pi) +
+ ftdholdpflfa pif + U^l d Pi 3 )s{p2 pi) 3 .
By the thermodynamic identity dw = T ds + V dp we have for the derivatives
(dzv/ds)p = T, {dwldp) s = V.
Hence
W2 — »i = Ti(s2si)+Vi(p2pi) +
+ wvidpdfa pi) 2 + \{&v\dpf) s (p>2 pif.
The volume Vz need be expanded only with respect top2—pi, since the second
term of equation (82.9) already contains the small difference p2 —pi, and an
expansion with respect to $2 — si would give a term of the form ($2 — si)(p2 —pi)>
which is of no interest. Thus
VtVi = (dVldpdsipzpj+^V/dp^fopiY.
Substituting this expansion in (82.9), we obtain
1 I dW \
^mm)*** (83,1)
Thus the discontinuity of entropy in a weak shock wave is of the third order
of smallness relative to the discontinuity of pressure.
In all cases that have been investigated, the compressibility —(dVjdp) s
decreases with increasing pressure, i.e. the second derivative
(dW\dp*) s > 0. (83.2)
It should be emphasised, however, that this is not a thermodynamic relation,
and cannot be derived by thermodynamic arguments. It is therefore possible
in principle that the derivative might be negative. We shall find several times
in what follows that the sign of the derivative (d 2 VI8p 2 ) s is very important
in gas dynamics. In future we shall assume it to be positive!
Let us draw through the point 1 (pi, Vi) in the pVplane two curves, the
shock adiabatic and the Poisson adiabatic. The equation of the latter is
s 2 — si = 0. By comparing this with the equation (83.1) of the shock adiabatic
near the point 1, we see that the two curves have contact of the second order at
this point, both the first and the second derivatives being equal. In order to
decide the relative position of the two curves near the point 1, we use the
fact that, according to (83.1) and (83.2), we must have s 2 > si on the shock
adiabatic for p2 > pi, while on the Poisson adiabatic S2 = *h The abscissa
of a point on the shock adiabatic must therefore exceed that of a point on
t For a perfect gas (0*F/fip*)g = (y+l)^/v 2 / >2 * This expression can be most simply obtained by
differentiating Poisson's adiabatic equation pVY = constant.
324
Shock Waves
§83
tfye Poisson adiabatic having the same ordinate p^. This follows at once
from the fact that, by the wellknown thermodynamic formula (dV\ds) v
— {Tjc v )(dVjdT) v , the entropy increases with the volume at constant
pressure for all substances which expand on heating, i.e. which have {dV\dT) v
positive. We can similarly deduce that, for p2 < pi, the abscissa of a point
on the Poisson adiabatic exceeds that of the corresponding point on the shock
adiabatic. Thus, near the point of contact, the two curves lie as shown in
Fig. 45 (HH' being the shock adiabatic and PP' the Poisson adiabatic)f,
both being concave upwards, by (83.2).
Fig. 45
For small p2pi and V2V1, formula (82.6) can be written, in the first
approximation, as j 2 = —{dpjdV) s (we take the derivative for constant
entropy, since the tangents to the two adiabatics at the point 1 coincide).
The velocities v\ and V2 are, in the same approximation, equal :
Vl = v 2 = v = jV = V[ V2(dp/dV) s ] = V( 8 P/dp)s.
This is just the velocity of sound c. Thus the rate of propagation of weak
shock waves is, in the first approximation, the velocity of sound :
v = c. (83.3)
From the properties of the shock adiabatic near the point 1 derived above
we can deduce a number of important consequences. Since we must have
$2 > $i in a shock wave, it follows Xh.2A.p2 > pi, i.e. the point 2 (j>2, V2) must
lie above the point 1. Moreover, since the chord 12 has a greater slope than
the tangent to the adiabatic at the point 1 (Fig. 43), and the slope of the
tangent is equal to the derivative (dp/dVi) Sl , we have p > (dpjdVi) Sl .
Multiplying both sides of this inequality by Vi 2 , we find
j2 Vl 2 = Vl 2 > _ VftdpldVi)* = (dpjd pi ) Sl = Cl 2 ,
where c\ is the velocity of sound corresponding to the point 1. Thus v\ > c\.
t If (dVjdT)p is negative, the relative position is reversed.
§84 The direction of "variation of quantities in a shock wave 325
Finally, from the fact that the chord 12 has a smaller slope than the tangent
at the point 2, it follows in like manner that v 2 < c 2 .\
§84. The direction of variation of quantities in a shock wave
The results of §83 show that, if the derivative (d 2 Vldp 2 ) s is assumed
positive, it can be demonstrated very simply that for weak shocks the con
dition of increasing entropy (s 2 > $i) necessarily means that
p 2 > pi, (84.1)
vi > c\, v 2 < C2. (84.2)
From the remark made concerning (82.6) it follows that, if p2 > pi, then
Vi > V 2 , (84.3)
and, since v±jVi = v 2 jV 2 = j, also
©i > v 2 . (84.4)
We shall now show that all these inequalities actually hold (still on the
assumption that (d 2 Vldp 2 ) s is positive) for shock waves of any intensity.
We shall therefore conclude, in particular, that, when gas passes through a
shock wave, it is compressed, the pressure and density increasing (E. Jouguet,
1904; G. ZemplEn, 1905)4 This means, graphically, that only the upper
branch of the shock adiabatic (above the point 1) has any real significance;
shock waves corresponding to points on the lower branch cannot exist. We
may also mention the following important result which can be derived from
the inequalities (84.2). Since a shock wave moves relative to the gas in front
of it with a velocity vi > c\, it is clear that no perturbation starting from
the shock wave can penetrate into that gas. In other words, the presence of
the shock has no effect on the state of the gas in front of it.
We shall now prove these statements, beginning with a preliminary
calculation. We differentiate the relations (82.5) and (82.8) with respect
to the quantities pertaining to gas 2, assuming the state of gas 1 to be un
changed. This means that pi, V± and wi are regarded as constants, while
pi, V 2 , W2 and also^ (which depends on £2 and V 2 ) are differentiated. From
(82.5) we obtain
Fid(j2) = dp 2 +j*dV 2 +V 2 d(P),
or
dp 2 +j 2 dV 2 = ( Vi  V 2 )d(j 2 ), (84.5)
t It can easily be shown in the same way that, when the derivative (8 2 V/dp 2 )s is negative, the con
dition s 2 > j x for weak shock waves implies that p 2 < /> x , while the velocities again satisfy v t > c lt
v 2 < c 2 .
{ If we change to a coordinate system in which gas 1 (in front of the shock wave) is at rest, and
the shock is moving, then the inequality v t > v 2 means that the gas behind the shock wave moves
(with velocity v t —v 2 ) in the same direction as the shock itself.
326 Shock Waves
and from (82.8)
dw 2 +p V 2 dV 2 = Wi 2  V 2 2 )d{j 2 ) y
or, expanding the differential dw 2 ,
T 2 ds 2 + V 2 dp 2 +j 2 V 2 dV 2 = i(V^ V 2 2 )d(j 2 ).
Substituting this equation in (84.5), we obtain
T 2 ds 2 = i(V 1 V 2 fd(P).
Hence we see that
d(j2)/d* 2 > 0,
i.e. j 2 increases with s 2 .
§84
(84.6)
(84.7)
>
Fig. 46
We now show that there can be no point on the shock adiabatic at which it
touches any line drawn from the point 1 (such as the point O is in Fig. 46).
At such a point the slope of the chord from the point 1 is a minimum, and/ 2
has a corresponding maximum, so that d(j 2 )jdp 2 = 0. We see from (84.6)
that in this case we also have ds 2 Jdp 2 = 0. Next, substituting in (84.5) the
differential dV 2 in the form dV 2 = (dV 2 fdp 2 ) Si dp 2 + (dV 2 lds 2 ) P2 ds 2 and ds 2
in the form given by (84.6), and dividing by dp 2 , we obtain
\ fy2 Is, J \ T 2 \ 8s 2 ) J dp 2
Hence it follows that, for d(j 2 )/dp2 = 0, we must have
Wf{W 2 \dp 2 ) Sl = \v 2 2 \c 2 2 = 0,
i.e. v 2 — c%\ conversely, if v 2 = c 2 , it follows that d(J 2 )jdp 2 = O.f
Thus, of the three equations
d(p)fdp2 = 0, ds 2 /dp 2 = 0, v 2 = c 2 , (84.8)
each implies the other two and all three would hold at the point O (Fig. 46).
t The expression in the braces can vanish only by chance, and this possibility is therefore
unlikely.
§84 The direction of variation of quantities in a shock wave 327
Finally, we have for the derivative of j 2 (dV2/ dp2)s 2 = — z>2 2 /f2 2 at the point O
dp 2 \ c 2 2 J J \ dp 2 */ S2 '
On account of the assumption that (d 2 Vjdp 2 ) s is positive, we therefore have
at O
d(v2lc 2 )ldp 2 < 0. (84.9)
It is now easy to show that such a point cannot exist on the shock adiabatic.
At points just above the point 1, ^2/^2 < 1 (see the end of §83). The equation
V2JC2 = 1 can therefore be satisfied only by an increase in ^2/^2 ; that is, at O
we should necessarily have d(v2/c2)ldp2 > 0, whereas by (84.9) the converse
is true. In an entirely similar manner, we can show that the ratio ^2/^2 also
cannot become equal to unity on the part of the shock adiabatic below the
point 1.
From the impossibility of the existence of a point such as O, which has
just been demonstrated, we can at once deduce from the graph of the shock
adiabatic that the slope of the chord from the point 1 (/>i, Vi) to the point 2
(p2, V2) decreases as we move up the curve, and/ 2 correspondingly increases.
From this property of the shock adiabatic and the inequality (84.7) it follows
immediately that the necessary condition S2 > si implies that P2 > P\ also.
It is also easy to see that, on the upper part of the shock adiabatic, the
inequalities V2 < C2, v\ > c\ hold. The former follows at once from the fact
that it holds near the point 1, and the ratio ^2/^2 can never become equal to
unity. The second inequality follows from the fact that every chord from the
point 1 to a point 2 above it is steeper than the tangent to the adiabatic at the
point 1, since the curve cannot behave as shown in Fig. 46.
The condition $2 > s\ and all three inequalities (84.1), (84.2) are therefore
satisfied on the upper part of the shock adiabatic. On the lower part, however,
none of these conditions holds. They are consequently equivalent, and if one
is satisfied so are all the others.
In the preceding discussion we have everywhere assumed that the deriva
tive (d 2 Vjdp 2 ) s is positive. If this derivative could change sign, it would no
longer be possible to draw from the necessity of $2 > *i any general conclu
sions concerning inequalities for the other quantities. It is important,
however, that the inequalities (84.2) for the velocities can be obtained by
quite different arguments, which show that shock waves in which those
inequalities do not hold cannot exist, even if their existence would not
be disproved by the purely thermodynamic arguments given above.
The reason is that we have still to discuss the subject of he stability of shock
waves. Let us suppose that a shock wave at rest is subjected to an infinite
simal displacement in a direction (say) perpendicular to its plane. It can be
shown that the result of such a displacement is that the shock wave is con
tinually accelerated in some direction, and it is clear that this demonstrates
the absolute instability of such a wave and the impossibility of its existence.
328
Shock Waves
§84
The displacement of the shock wave is accompanied by infinitesimal
perturbations in the gas pressure, velocity, etc. on both sides of the surface
of discontinuity. These perturbations near the shock are then propagated
away from it with the velocity of sound (relative to the gas) ; this, however,
does not apply to the perturbation in the entropy, which is transmitted only
with the gas itself. Thus an arbitrary perturbation of the type in question can
be regarded as consisting of sound disturbances propagated in gases 1 and 2
on both sides of the shock wave, and a perturbation of the entropy; the latter,
which moves with the gas, will evidently occur only in gas 2 behind the shock.
In each of the sound disturbances, the changes in the various quantities are
related by certain formulae which follow from the equations of motion (as
in any sound wave, §63), and therefore any such disturbance is specified
by only one parameter.
1'1>C, / 2 <C 2
"l< C l V Z< C 2
Fig. 47
Let us now compute the number of possible sound disturbances. It
depends on the relative magnitudes of the gas velocities v\, vi and the sound
velocities c\, c^. We take the direction of motion of the gas (from 1 to 2)
as the positive direction of the xaxis. The rate of propagation of the distur
bance in gas 1 relative to the stationary shock wave is u± = v±±ci, and in
gas 2 it is 14,2 = ^2 ± £2 Since these disturbances must be propagated away from
the shock wave, it follows that u\ < 0, m > 0.
Let us suppose that v\> c\, vi < c^. Then it is clear that both values
u\ = v\± c\ are positive, while only v% + c% of the two values of ui is positive.
This means that the sound disturbances in which we are interested cannot
exist in gas 1, while in gas 2 there can be only one, which is propagated relative
to the gas with velocity c^. The calculation in other cases is similar.
The result is shown in Fig. 47, where each arrow corresponds to one sound
disturbance, propagated relative to the gas in the direction shown by the
arrow. Each sound disturbance is defined, as stated above, by one parameter.
Furthermore, in all four cases there are two other parameters, one determining
the entropy perturbation propagated in gas 2 and one determining the dis
placement of the shock wave. For each of the four cases in Fig. 47, the
§85 Shock waves in a perfect gas 329
number in a circle shows the total number of parameters, thus obtained,
which define an arbitrary perturbation arising from the displacement of the
shock wave.
The number of boundary conditions which must be satisfied by a pertur
bation on the surface of discontinuity is three (the continuity of the mass,
energy and momentum fluxes). The solution of the stability problem is
effected by prescribing the displacement of the shock wave (and therefore the
perturbations in all the other quantities) in a form proportional to e nt , and
determining the possible values of Q by means of the boundary conditions;
the existence of real positive values of Q indicates absolute instability. In
all except the first of the cases shown in Fig. 47, the number of parameters
available exceeds the number of equations given by the boundary conditions
at the discontinuity. In these cases, therefore, the boundary conditions
admit any (and therefore any positive) value of O, and the shock wave is
absolutely unstable. In the one case v\ > c\, V2 < c%, however, the number of
parameters just equals the number of equations, and these therefore give a
definite value of Q.. It is evident, without writing down the equations, that
this value must be Q, = 0, since the problem contains no parameter of the
dimensions sec 1 which could determine a value of Q. different from zero but
not arbitrary. There is therefore no such instability in this case.
Thus we see that the inequalities (84.2) for the velocity of the shock
wave are necessary for the shock to exist, whatever the thermodynamic
properties of the gas.
In order to decide the stability of shock waves for which the condition
(84.2) is satisfied, we should have to investigate also the other possible modes
of instability. One of these is instability with respect to perturbations of the
kind considered in §30 (characterised by periodicity in the direction parallel
to the surface of discontinuity and forming "ripples" on this surface). We
shall not perform the calculations here, but merely mention that shock waves
are almost always stable with respect to such perturbations. Instability can
occur only for certain very special forms of the shock adiabatic, which seem
hardly ever to occur in Nature; they all require that the derivative (d 2 Vjdp 2 ) 8
should be of variable sign. j
A shock wave might also, in principle, be unstable with respect to break
up into more than one surface of discontinuity. This problem has not been
adequately investigated, but such instabilities may likewise occur only for
certain very special types of shock adiabatic.
§85. Shock waves in a perfect gas
Let us apply the general relations obtained in the previous sections to
shock waves in a perfect gas. The heat function of a perfect gas is given by
the simple formula w = ypVI(y—l). Substituting this expression in (82.9),
■f See S. P. D'yakov, Zhurnal eksperimental'noi i teoreticheskoi fiziki 27, 288, 1954; V. M.
Kontorovich, ibid. 33, 1525, 1957; Soviet Physics JETP 6 (33), 1179, 1958.
330
Shock Waves
§85
we have after a simple transformation
Yl_ = (r+l)j>i + (yl)j>2
V ± (yl)p 1 + (y+l)p 2 '
(85.1)
Using this formula, we can determine any of the quantities p h V ly p 2 , V 2
from the other three. The ratio V 2 \V\ is a monotonically decreasing function
of the ratio p 2 /p h tending to the finite limit (y  l)/(y + 1). The curve showing
p 2 as a function of V 2 for given p h V\ (the shock adiabatic) is represented in
Fig. 48. It is a rectangular hyperbola with asymptotes V 2 \V\ = (y l)/(y + 1),
pz/pi =  (y l)/(y + 1). As we know, only the upper part of the curve, above
the point V 2 jVi = p 2 \p\ — 1, has any real significance; it is shown in Fig. 48
(for y — 1 4) by a continuous line.
For the ratio of the temperatures on the two sides of the discontinuity
we find, from the equation of state for a perfect gas T 2 \T\ = p 2 V 2 \p\V\ y
that
72
7\
Pi (y+l)/>i + (yl)/>2
(85.2)
p! (yl)pi + (y+l)p 2
For the flux density; we obtain from (82.6) and (85.1)
i 2 = {(y l)/>i + (y + l)fr}/2Fi, (85.3)
and then for the velocities of propagation of the shock wave relative to the
gas before and behind it
^i 2 = ^i{(yi)/>i+(y+l)M
*2 2 = Wi(v+ l)/»i + (y lW 2 /{(y l)/>i+(y+ l)/> 2 }.
(85.4)
§85 Shock waves in a perfect gas 331
We may derive limiting results for very strong shock waves, in which p%
is very large compared with^i.f From (85.1) and (85.2) we have
v 2 lv 1 = P1 i P 2 = (yi)i( y +i), ra/ri = (yiMy+i)fr ( 85  5 )
The ratio T%\T\ increases to infinity with pzjpi, i.e. the temperature discon
tinuity in a shock wave, like the pressure discontinuity, can be arbitrarily
great. The density ratio, however, tends to a constant limit; e.g., for a
monatomic gas the limit is pi — 4/>i, and for a diatomic gas p2 = 6pi. The
velocities of propagation of a strong shock wave are
oi = VWy+ifaVi}* V* = V(i(y i)W/(y + 1» (85.6)
They increase as the square root of the pressure p^.
Finally, we may give some formulae useful in applications, which express
the ratios of densities, pressures and temperatures in a shock wave in terms of
the Mach number Mi = v\\c\. These formulae are easily derived from the
foregoing results :
P2IP1 = t*/t* = (y+ l)Mi*/{(y l)Mi* + 2}, (85.7)
p 2 / Pl = 2yMi2/(y+ 1) (y  l)/(y + 1), (85.8)
Ta/Ti = {2yM 1 2( y  l)}{(y l)Mi2 + 2}/(y+ l)2Mi. (85.9)
The Mach number M2 is given in terms of Mi by
M 2 2 = (2 + (y~l)Mi2}/{2yMi2(yl)}. (85.10)
PROBLEMS
Problem 1. Derive the formula v x v 2 = c* 2 , where c* is the critical velocity.
Solution. Since to+^v 2 is continuous at a shock wave, we can define a critical velocity
which is the same for gases 1 and 2 by
yp 1 .12 y P 2 ,i2 r+1 2
* >
(yl)pi (y~l)p2 2(yl)
cf. (80.7). Determining p 2 lp 2 and Pifpi from these equations and substituting in
p2 p\
Vl — V* =
P2V2 piVl
(obtained by combining (82.1) and (82.2)), we obtain
y+1 / r.2
(t>ii*)(l — ) =0.
2y \ V1V2 
Since v t # v 2 , this gives the required relation.
Problem 2. Determine the value of the ratio p 2 /Pi, for given temperatures T u T 2 at a dis
continuity in a perfect gas with a variable specific heat.
t It is necessary that not only p 2 ^> p x but p 2 ^> (y+ l)/>i/(y — !)•
332 Shock Waves §85
Solution. In the general case of a perfect gas with variable specific heat, we can say only
that u) (like e) is a function of temperature alone, and that p, Fand Tare related by the equa
tion of state pV = RTIvl. Solving equation (82.9) for pjp u we obtain
Pi RT ' 2ft WU i?ft 2ft J + ft /'
where Wi =* w(2 , 1 ), w 2 = w(T 2 ).
Problem 3 . A plane sound wave meets normally a shock wave in a perfect gas. Determine the
intensity of sound transmitted by the shock wave (D. I. Blokhintsev, 1945). f
Solution. Since a shock wave is propagated with supersonic velocity relative to the gas
in front of it, no sound wave can be reflected from it. In gas 2, behind the discontinuity, an
ordinary lsentropic transmitted sound wave is propagated, and also a perturbation of the
entropy (at constant pressure), which is propagated with the moving gas itself.
We consider the process in a coordinate system in which the shock wave is at rest, and the
gas moves through it in the positive direction of the a;axis, the incident sound wave being
propagated in this direction also. The perturbations on the two sides of the discontinuity are
related by conditions obtained by varying the boundary conditions (82.1)(82.3). As a result
of the sound disturbance, the shock wave also begins to oscillate; denoting its oscillatory
velocity by 8u, we must write the change in the velocities v u v 2 in the boundary conditions
as 8vj_ — 8u, 8v 2 — 8u. Thus J
ViSpi+pxfiviSll) = V 2 8p 2 +p2(8v28u),
Spi+vi 2 8p 1 + 2p 1 v 1 (8v 1 8u) = 8p2+v2 2 8p2 + 2p 2 v z {8v28u),
8w 1 +v 1 (8vi8u) = 8zv 2 +V2(8v28u).
In the incident sound wave we have
$Si = 0, S^i = (Ci/p{)8pi = 8pi/dpi, 8W! = 8pil P1 .
The perturbation in medium 2 is composed of the sound wave and the "entropy wave",
which we denote by one and two primes respectively:
S* 2 ' =* 0, 8V = (C2/p2)8p2 = 8p 2 '/c 2 p2, 8w 2 = 8p2'[p 2 ,
8p2 r ' = 0, 8v 2 " = 0, 8zv 2 " = T28S2" = C2*8 P2 "lp2{y\)
(for a perfect gas (dsjdp)^ — —cjp).
These relations enable us to express all quantities in the transmitted waves in terms of the
corresponding quantities in the incident wave. The ratio of pressures in the sound waves is
found to be
8p 2 ' Mi + 1 /2( y l)M 1 M 2 2 (Mi2l)(Mifl)[( y l)Mi2 + 2]
8p! M 2 + 1 I 2( r  l)M 2 2(Mi2  1)  (M 2 + l)[( y  l)Mi2 + 2]
For a weak shock wave (p 2 —pi <^Pi) we find
8P2 . y+1 p2~Pl
« 1 H ,
8pl 2y />!
f The solution of the more general problem of oblique incidence of sound on a shock wave in an
arbitrary medium is given by V. M. Kontorovich, Zhurnal experimental' noi i teoreticheskoi fiziki
33, 1527, 1957; Soviet Physics JETP 6 (33), 1180, 1958.
J We here denote the variable parts of quantities by 8 instead of the usual prime.
§86
Oblique shock waves
333
and in the opposite limiting case of a strong shock wave
8p2 1 p2
SpT ~ y+V[2y(yl)] pi
In both cases the pressure amplitude in the transmitted wave is greater than that in the
incident wave.
§86. Oblique shock waves
Let us consider a steady shock wave, and abandon the system of co
ordinates used hitherto, in which the gas velocity is perpendicular to the shock
surface element considered. The streamlines can intersect the surface of
such a shock wave at any angle,f and in doing so are "refracted" : the tan
gential component of the gas velocity is unchanged, while the normal com
ponent is, according to (84.4), diminished: v\ t = ®2t, ^i» > ^2». It is there
fore clear that the streamlines "approach" the shock wave as they pass through
it (cf. Fig. 49). Thus the streamlines are always refracted in a definite direc
tion in passing through a shock wave.
Fig. 49
The motion behind a shock wave may be either subsonic or supersonic
(only the normal velocity component need be less than the velocity of sound
C2) ; the motion in front of it is necessarily supersonic. If the gas flow on
both sides is supersonic, every disturbance must be propagated along the
surface in the direction of the tangential component of the gas velocity. In
this sense we can speak of the "direction" of a shock wave, and distinguish
shock waves leaving and reaching any point (as we did for characteristics, the
motion near which is always supersonic; see §79). If the motion behind the
shock is subsonic, there is strictly no meaning in speaking of its "direction",
since perturbations can be propagated in all directions on its surface.
We shall derive a relation between the two components of the gas velocity
after it has passed through an oblique shock wave, supposing that we have a
perfect gas. We take the direction of the gas velocity vi in front of the shock
as the #axis; let <f> be the angle between the shock and the #axis (Fig. 49).
f The only restriction is that the normal velocity component v ln exceeds c x .
334
Shock Waves
§86
The continuity of the velocity component tangential to the shock means that
vi cos <f> *= V2x cos <f> + V2y sin <f>, or
tan<£ = (viV2 X )jvzy.
(86.1)
Next we use formula (85.7), in which v\ and v 2 denote the velocity components
normal to the plane of the shock wave and must be replaced by vi sin (f> and
V2x sin <f> — vzy cos <£, so that
V2x sin^ — V2y cos<f> y—\
v\ sin <f>
+
2d 2
y+1 (y+ l)^i 2 sin 2 </>
(86.2)
Fig. 50
We can eliminate the angle § from these two relations. After some simple
transformations, we obtain the following formula which determines the re
lation between v 2x and v 2y (for given vi and c{) :
V 2 y 2 = (V 1 V 2X ) 2
2(V 1 C 1 *lv 1 )l(y +l )fo i  V 2x )
Vi  V2 X + 2ci 2 /(y + 1 )V!
(86.3)
This formula can be more intelligibly written by introducing the critical
velocity. According to Bernoulli's equation and the definition of the critical
velocity, we have m + W = ^i 2 + ci 2 /(yl) = (y+l>v72( y l) (see §85,
Problem 1), whence
c**= [(yl)^ 2 + 2 tl 2 ]/( r +l).
Using this in (86.3), we obtain
V 2 y 2 = (VlV2x) 2 ~
VlV2xC* 2
(86.4)
(86.5)
2vi 2 /(y + 1 )  v±V2x + C* 2
Equation (86.5) is called the equation of the shock polar. Fig. 50 shows a
praph of the function V2y{v2x)\ it is a cubic curve, called a strophoid. It
crosses the axis of abscissae at the points P and Q, corresponding to V2x
§86 Oblique shock waves 335
= c^\v\ and v% x = »i«t A li ne (OS in Fig. 50) drawn from the origin at an
angle x to the axis of abscissae gives, by the length of the segment between O
and the point where it intersects the shock polar, the gas velocity behind a
discontinuity which turns the stream through an angle x There are two such
intersections (A, B), i.e. two different shock waves correspond to a given
value of X The direction of the shock wave also can be immediately deter
mined from the shock polar: it is given by the direction of the perpendicular
from the origin to the line QB or QA (Fig. 50 shows the angle for a shock
corresponding to the point B). As % decreases, the point A approaches P,
corresponding to a normal shock {<f> = \tt) with V2 = c^\v\. The point B
approaches Q\ the intensity of the shock (velocity discontinuity) tends
to zero, and the angle cf> tends, as it should, to the Mach angle a =
sin 1 (cijvi)\ the tangent to the shock polar at Q makes an angle £77+ a with
the axis of abscissae.
From the shock polar we can immediately derive the important result that
the angle of deviation x of the stream at the shock wave cannot exceed a certain
maximum value xmax, corresponding to the tangent from O to the curve.
This quantity is, of course, a function of the Mach number Mi = vijci,
but we shall not give the expression for it, which is very cumbersome. For
Mi = 1, xmax = 0; as Mi increases, xmax increases monotonically, and tends
to a finite limit as Mi >• 00. It is easy to discuss the two limiting cases.
If the velocity v\ is near to c+, then V2 is so also, and the angle x is small;
the equation (86.5) of the shock polar can then be written in the approximate
formj
x 2 = (y+l)(v 1 v 2 ) 2 (v 1 +V22c m )l2cJ, (86.6)
where we have put v% x ~ ^2, V2y ~ c*x i n view of the smallness of x Hence
we easily findff
W(y +l) ("i ,\« V2 , „ n , , s , 7 ,
^WJW. 1 ) = 3V3(y+D (Ml1)  (86 ' 7)
In the opposite limiting case Mi = 00 (i.e. Mi* = V[(y+l)/(7 — l)])> tne
shock polar degenerates to a circle which meets the axis of abscissae at the
points c^[(y— l)/(y + 1)] and c^y/fty + l)/(y — 1)]. It is easy to see that we
then have
Xmax = sini(l/ y ); (86.8)
f The strophoid actually continues in two branches from the point v 2 x = v x (which is a double
point) to infinite f 2 j/> these are not shown in Fig. 50. They have a common asymptote
vzx = c* 2 lvi + 2vi((y+l).
The points on these branches have no physical significance; they would give values for v 2 x and V2y
such that v 2 Jv ln > 1, which is impossible.
J It is easily seen that equation (86.6) holds also for any (nonperfect) gas, provided that (y+1)
is replaced by 2a* (95.2).
If It may be noted that this dependence of Xmax on M x — 1 is in agreement with the general simi
larity law (118.7) for transonic flow.
336
Shock Waves
§86
for air this is 456°. Figure 51 shows a graph of xmax as a function of Mi for
air; the Upper curve is a similar graph for flow past a cone (see §105).
The circle v% = c* cuts the axis of abscissae between the points P and
Q (Fig 50), and therefore divides the shock polar into two parts correspond
ing to subsonic and supersonic gas velocities behind the discontinuity. The
point where this circle crosses the polar lies to the right of, but very close to,
the point C; the whole segment PC therefore corresponds to transitions to
subsonic velocities, while CQ (except for a very small segment near C)
corresponds to transitions to supersonic velocities.
Fig. 51
For given Mi and </>, the pressure change in the shock wave is given by
p 2 2yMi 2 sin 2 <i(yl)
T = 1 /n ; < 86  9 )
Pi (y+i)
this is formula (85.8) with Misin<£ in place of Mi. This ratio increases
monotonically when the angle <f> increases from its smallest value sin 1 (l/Mi)
(when P2IP1 = 1) to \n, i.e. as we move along the shock polar from Q to P.
The two shock waves determined by the shock polar for a given deviation
angle x are often said to belong to the weak and strong families. A shock
wave of the strong family (the segment PC of the polar) is strong (the ratio
P2IP1 is large), makes a large angle <f> with the direction of the velocity vi,
and converts the flow from supersonic to subsonic. A shock wave of the weak
family (the segment QC) is weak, is inclined at a smaller angle to the stream,
and almost always leaves the flow supersonic.
PROBLEMS
Problem 1 . Derive the formula giving the angle of deviation x of the velocity in an oblique
shock wave (in a perfect gas) in terms of Mi = vjc^ and the angle ^ between the shock
wave and the direction of the velocity v x (Fig. 49) :
cotx = tan^
.2(Mi2sin2«il) J
2(M r
§87 The thickness of shock waves 337
Problem 2. Derive the formula giving the number M 2 = vjc 2 in terms of M x and <£:
2 + (yl)Mi 2 2Mi 2 cos 2 <£
M2 2 = , .  . — +
2yMi 2 sin 2 <f>  (y  1) 2 + (y  l)Mi 2 sin 2 <£
§87. The thickness of shock waves
Hitherto we have regarded shock waves as geometrical surfaces of zero
thickness. We shall now consider the structure of actual surfaces of dis
continuity, and we shall see that shock waves in which the discontinuities
are small are in reality transition layers of finite thickness, the thickness
diminishing as the magnitude of the discontinuities increases. If the dis
continuities are not small, the change occurs so sharply that the concept of
thickness is meaningless.
To determine the structure and thickness of the transition layer we must
take account of the viscosity and thermal conductivity of the gas, which we
have hitherto neglected.
The relations (82.1)(82.3) for a shock wave were obtained from the
constancy of the fluxes of mass, momentum and energy. If we consider a
surface of discontinuity as a layer of finite thickness, these conditions must be
written, not as the equality of the quantities concerned on the two sides of
the discontinuity, but as their constancy throughout the thickness of the
layer. The first condition, (82.1), is unchanged:
pv =j= constant. (87.1)
In the other two conditions additional fluxes of momentum and energy, due
to internal friction and thermal conduction, must be taken into account.
The momentum flux density (in the xdirection) due to internal friction is
given by the component — & ' xx of the viscosity stress tensor ; according to the
general expression (15.3) for this tensor, we have a xx — ON + Qjdvjdx.
The condition (82.2) then becomes
p+pv 2 — (jft + Qdv/dx = constant.
As in §82, we introduce the specific volume V in place of the velocity v = jV.
Since j = constant, dvjdx = jdVjdx, so that
P +j 2 V— (f 77 + £)/ dVjdx = constant.
At great distances from the shock wave, the thermodynamic quantities are
constants, i.e. they are independent of x; in particular, dVjdx = 0. We
denote by a suffix 1 the values of quantities far in front of the shock wave.
Then we can put the constant equal to Pi+j 2 Vi, obtaining
ppi+P(V V^^rj + QjdV/dx = 0. (87.2)
Next, the energy flux density due to thermal conduction is — k dTjdx. That
due to internal friction is —a' X iVi or, since the velocity is along the #axis,
— o'xxv = —(£rj+£)vdv/dx. Thus the condition (82.3) can be written
pv{w\\v 2 ) — {%q + Qv dvjdx— k dTjdx = constant.
338 Shock Waves §87
Again putting v = jV, we can obtain the final form
v>+$pV*j($r} + QVdV/6x(Klj)dT/dx = wi + i/W (87.3)
We shall here consider shock waves in which all the discontinuities are
small. Then all the differences V— Vi, p—pi, etc. between the values inside
and outside the transition layer are also small. In (87.2) we expand VV\
in powers oip—pi and s — si, taking the pressure and entropy as the indepen
dent variables. It is seen from the relations obtained below that 1/S (where
S is the thickness of the discontinuity) is of the first order in p —pi, and the
difference s — s± is of the second order.f Hence we can write, neglecting quan
tities of the third order,
FFi = (dVldp) s {pp 1 )+^8Wl8p^ s {pp 1 f + (8V/8s) p (ss 1 ).
The values of all the coefficients are, of course, taken outside the transition
layer (i.e. for p = p h s = s{). Substituting this expansion in (87.2), we
obtain
[1 + (dV/8p) s p](p  Pl ) + ij2^2 V/8p 2 )s (p _ pi) 2 + (8V/8s) p (s  S1 )P
= (tn + QjdV/dx.
The derivative dF/dx can be written
dV _ / 8V \ dp / 8V\ ds
dx \ dp J s dx \ 8s Jp dx
Differentiation with respect to x increases the order of smallness by one,
since 1/8 is of the first order; the derivative dp/dx is therefore of the second
order, and ds/dx of the third order. The term in ds/dx can therefore be omit
ted. Thus the condition (87.2) becomes
[l^Vl^splipp^HPi^V/Sp^ipp^ + idV/Ss^iss^P
= (ft + Z>){Wl8p) s (dp/dx) j. (87.4)
Next, we multiply each term of (87.2) by \{V+ V\) and subtract from
equation (87.3). The result is
itot 0l )\{pp 1 yy+ VjiMn+ZW v 1 )^ K  ^ =0.
dx j ax
The third term, which contains the product (V—V\) dVjdx, is of the third
order, and may be omitted :
(wwi)i(p^)i)(F+ri)(/c/;")dT/dx = 0.
The first two terms are just the expression which we expanded in powers of
p—pi smds—si in deriving formula (83.1). The first and second order terms
in p —p\ in this expansion are zero, and we have as far as terms of the second
t The total entropy discontinuity s 2 — s x is, as we have seen in §83, of the third order relative to
the pressure discontinuity p 2 —pi> whereas s— Sj is of only the second order in p—p x . The reason is
that, as we shall show below, the pressure in the transition layer varies monotonically from p x to p 2 ,
whereas the entropy does not vary monotonically; it has a maximum within the layer.
§87 The thickness of shock waves 339
order just T(ss{). The derivative dTjdx can be written
dT _ / dT\ dp l 8T\ ds ^ / dT \ dp
dx \ dp J s dx \ ds / „ dx \ dp / s dx
The result is
""^jUlsr (87  5)
Substituting this expression for s — si in (87.4), we find
"■(J),» w " + [' t (f)/] < ' w
(y(?l(f ).*;»©.)&■ <«">
The flux j is, in the first approximation, j = vjV « c/F (see (83.3)).
This expression can be substituted on the righthand side of (87.6), but it
will not serve on the lefthand side; further terms have to be included
in^ 2 . These terms could be obtained from (87.2), for instance. It is simpler,
however, to argue as follows. At great distances on both sides of the surface
of discontinuity, the righthand side of (87.6) is zero, since dpjdx is zero.
At such distances the pressure is p\ or />2 That is, we can say that the
quadratic in p on the left of (87.6) has the zeros p\ and p2. By a wellknown
theorem of algebra, it can therefore be written as the product (p —pi)(p —p?)
multiplied by the coefficient of p 2 , ^j 2 (d 2 Vldp 2 ) s .
Thus we have the following differential equation for the function p(x) :f
\ I dW\ V 3 { k I dT \ / 8V \ n \ dp
From the thermodynamic formulae for derivatives, {dV\ds) v = (dTjdp) s ; it is
easy to see that the coefficient of — dp/dx on the righthand side of the above
equation is 2V 2 a, where a is related to the soundabsorption coefficient y
(77.6) by y = aco 2 . Thus
dp 1 1 d 2 V\
s4»^M.»^ w  (87  7)
Integration gives
\V 2 a r dp
x = —
J (P:
(d*V/dp 2 ) s J (p Pl )(pp 2 )
4aV2 p i(p 2+pi )
tanh 1 h constant.
i(p2pi)(d 2 V/dp 2 ) s &P2P1)
t In considering a weak shock wave we can regard the viscosity and the thermal conductivity as
constants.
340 Shock Waves §87
Putting the constant equal to zero, we have
P ~ KP2 +pi) = I(p2 pi) tanh(*/S), (87.8)
where
3 = 8aV 2 /(p 2 pi)(d 2 V/dp 2 ) s . (87.9)
This gives the manner of variation of the pressure between the values p\
and pi which it takes at great distances on the two sides of the shock wave.
The point x — corresponds to the median value of the pressure, i(j>i+P2)
For x >; ±oo, the pressure tends asymptotically to p\ and p2. Almost
the whole change from p\ to p2 occurs over a distance of the order of S,
which may be called the thickness of the shock wave. We see that this is the
less, the stronger the shock, i.e. the greater the pressure discontinuity.
The variation of the entropy across the discontinuity is obtained from
(87.5) and (87.8)
k i 8T\ i dW \ 1
J  Sl = T6^^(^) s (^i)^^^^P) ( 87  10 >
From this we see that the entropy does not vary monotonically, but has a
maximum inside the shock, at x = 0. For x = ± oo this formula gives
s = si in either case; this is because the total entropy change $2— si is of the
third order inp^—pi (cf. (83.1)), whereas s — si is of the second order.
Formula (87.8) is quantitatively valid only for sufficiently small differences
p2—pi We can, however, use (87.9) qualitatively to determine the order of
magnitude of the thickness in cases where the difference p2~Pi is of the
same order of magnitude as pi and p2 themselves. The velocity of sound in
the gas is of the same order as the thermal velocity v of the molecules. The
kinematic viscosity is, as we know from the kinetic theory of gases, v ~ h
~ Ic, where / is the mean free path of the molecules. Hence a ~ Ijc 2 ; an
estimate of the thermalconduction term gives the same result. Finally,
(d 2 Vjdp 2 ) s ~ Vjp 2 , and pV ~ c 2 . Using these relations in (87.9), we obtain
8 ~/. (87.11)
Thus the thickness of a strong shock is of the same order of magnitude as the
mean free path of the gas molecules.f In macroscopic gas dynamics, however,
where the gas is treated as a continuous medium, the mean free path must be
taken as zero. It follows that the methods of gas dynamics cannot strictly be
used alone to investigate the internal structure of strong shock waves.
A considerable increase in the thickness of a shock wave may be caused by
the presence in the gas of comparatively slow relaxation processes (slow
chemical reactions, a slow energy transfer between different degrees of free
dom of the molecule, and so on). This topic has been discussed by Ya. B.
Zel'dovich (1946).
t A strong shock wave causes a considerable increase in temperature; / denotes the mean free
path for some mean temperature of the gas in the shock.
§87
The thickness of shock waves
341
Let t be of the order of magnitude of the relaxation time. Both the
initial and the final states of the gas must be states of complete equilibrium;
it is therefore immediately clear that the total thickness of the shock wave will
be of the order of tz>i, the distance traversed by the gas in the time t. It
is also found that, if the shock strength is above a certain limit, its structure
becomes more complex; this may be seen as follows.
p 2
Fig. 52
In Fig. 52 the continuous curve shows the shock adiabatic drawn through a
given initial point 1, on the assumption that the final states of the gas are
states of complete equilibrium; the slope of the tangent at the point 1 gives
the "equilibrium" velocity of sound, denoted in §78 by cq. The dashed curve
shows the shock adiabatic through the same point 1, on the assumption that
the relaxation processes are "frozen" and do not occur. The slope of the
tangent to this curve at the point 1 gives the velocity of sound denoted in
§78 by <v
If the velocity of the shock wave is such that Co < v± < c^, the chord
12 lies as shown in Fig. 52 (the lower chord). In this case we have a simple
increase in the shock thickness, all intermediate states between the initial
state 1 and the final state 2 being represented in the pVplane by points on
the segment 12.)
If, however, v\> c^, the chord takes the position 11 '2'. No point lying
between 1 and 1' corresponds to any actual state of the gas ; the first real point
(after 1) is 1', which corresponds to a state in which the relaxation equili
brium is no different from that in state 1. The compression of the gas from
state 1 to state V occurs discontinuously, and afterwards (over distances
~ v\r) it is gradually compressed to the final state 2'
t This follows from the fact that (neglecting ordinary viscosity and thermal conduction) all the
states through which the gas passes satisfy the equations of conservation of mass, pv = j = constant,
and of momentum, p\pV = constant (cf. the similar but more detailed discussion in §121).
12
342 Shock Waves §88
§88. The isothermal discontinuity
The discussion of the structure of a shock wave in §87 involves the assump
tion that the viscosity and thermal conductivity are of the same order of
magnitude, as is usually the case. The case where x > v is a ls° possible,
however. If the temperature is sufficiently high, additional heat is transferred
by thermal radiation in equilibrium with the matter. Radiation has a much
smaller effect on the viscosity (i.e. the momentum transfer), and so v may
be small compared with x We shall now see that this inequality leads to a
very important difference in the structure of the shock wave.
Neglecting terms in the viscosity, we can write equations (87.2) and (87.3),
which determine the structure of the transition layer, as
p+PV = p 1 +jW lf (88.1)
J rp
 = w+ lj2 V 2_ Wl _lj2 Vl 2. ( 882 )
j dx
The righthand side of (88.2) is zero only at the boundaries of the layer.
Since the temperature behind the shock wave must be higher than that in
front of it, it follows that we have
dT/dx > (88.3)
everywhere in the transition layer, i.e. the temperature increases monotoni
cally.
All quantities in the layer are functions of a single variable, the coordinate
x, and therefore are functions of one another. Differentiating (88.1) with
respect to V, we obtain
\ dTJvdV \8VJt
The derivative (dpjdT) v is always positive in gases. The sign of the derivative
dTfdV is therefore the reverse of that of the sum (dpjdV)T +j 2  In state 1 we
have j 2 < —(dpidVil)s (since ^i > c{), and, since the adiabatic compressibility
is always less than the isothermal compressibility,; 2 > — {dp\jdVi)T. On side
1, therefore, dT\]dV\ < 0. If this derivative remains negative everywhere in
the transition layer, then, as the gas is compressed (V decreasing), the
temperature increases monotonically, in accordance with (88.3), from side 1
to side 2, In other words, we have a shock wave whose thickness is much
increased by the high thermal conductivity (possibly to such an extent that
even to call it a shock wave is mere convention).
If, however, the shock is so strong that
P < (8p 2 ldV 2 )T, (88.4)
then we have in state 2 d^/dF^ > 0, so that the function T{V) has a maxi
mum somewhere between V\ and Vi (Fig. 53). It is clear that the transition
§88
The isothermal discontinuity
343
from state 1 to state 2, with V changing continuously, then becomes im
possible, since the inequality (88.3) can not be satisfied everywhere.
Consequently, we have the following pattern of transition from the initial
state 1 to the final state 2. First comes a region where the gas is gradually
compressed from the specific volume V\ to some V (the value for which
T(V) = T 2 for the first time; see Fig. 53); the thickness of this region is
determined by the thermal conductivity, and may be considerable. The
compression from V to V 2 then occurs discontinuously, the temperature
remaining constant at T 2 . This may be called an isothermal discontinuity.
Fig. 53
Let us determine the variation of the pressure and density in an isothermal
discontinuity, assuming that we have a perfect gas. The condition of con
tinuity of momentum flux (88.1), applied to the two sides of the discontinuity,
gives p'+jW = p2+j 2 V 2 . For a perfect gas V = RTj^p; since T = T 2 ,
we have
, j 2 RT 2 pRT 2
P' + — — = P2 + J ——.
\xp fip 2
This quadratic equation for p' has the solutions p' = p 2 (trivial) and
p' =pRT 2 / H ,p2=pV 2 . (88.5)
We can express;* 2 in the form (82.6), obtaining p' = (p 2 pi)V 2 l(Vi V 2 \
and, substituting V 2 fVi from (85.1), we have
/»' = [(r+i)/»i+(yi)M (88.6)
Since we must have p 2 > p', we find that an isothermal discontinuity occurs
only when the ratio of the pressures p 2 and pi satisfies
P2/P1 > (y+l)/(3y) (88.7)
(Rayleigh 1910). This condition can, of course, be obtained directly from
(88.4).
344 Shock Waves §89
Since, for a given temperature, the gas density is proportional to the
pressure, the density ratio in an isothermal discontinuity is equal to the
pressure ratio;
P'lf* = V2IV = p'lpz. (88.8)
§89. Weak discontinuities
Besides surface discontinuities, at which the quantities p, p, v etc. are
discontinuous, we can also have surfaces at which these quantities, though
remaining continuous, are not regular functions of the coordinates. The
irregularity may be of various kinds. For example, the first spatial derivatives
of p, p, v etc. may be discontinuous on a surface, or these derivatives may
become infinite; or higher derivatives may behave in the same manner.
We call such surfaces weak discontinuities, in contrast to the strong discon
tinuities (shock waves and tangential discontinuities), in which the quantities
p, p, v, ... themselves are discontinuous.
It is easy to see from simple considerations that weak discontinuities are
propagated relative to the gas (on either side of the surface) with the velocity
of sound. For, since the functions />, p, v, ... themselves are continuous,
they can be "smoothed" by modifying them only near the surface of dis
continuity, and only by arbitrarily small amounts, in such a way that the
smoothed functions have no singularity. The true distribution of the
pressure, say, can thus be represented as a superposition of a perfectly smooth
function po, free from all singularities, and a very small perturbation p'
of this distribution near the surface of discontinuity; and the latter, like any
small perturbation, is propagated, relative to the gas, with the velocity of
sound.
It must be emphasised that, for a shock wave, the smoothed functions would
differ from the true ones by quantities which in general are not small, and
the foregoing arguments are therefore invalid. If, however, the discon
tinuities in the shock wave are sufficiently small, those arguments are again
applicable, and such a shock wave is propagated with the velocity of sound,
a result which was obtained by another method in §83.
If the flow is steady in a given coordinate system, then the surface of dis
continuity is at rest in that system, and the gas flows through it. The gas
velocity component normal to the surface must equal the velocity of sound.
If we denote by a the angle between the direction of the gas velocity and the
tangent plane to the surface, then v n = v sin a = c, or sin a = c\v, i.e. a
surface of weak discontinuity intersects the streamlines at the Mach angle.
In other words, a surface of weak discontinuity is one of the characteristic
surfaces, a result which is entirely reasonable if we recall the physical signi
ficance of the latter : they are surfaces along which small perturbations are
propagated (see §79). It is clear that, in steady flow of a gas, weak discon
tinuities can occur only at velocities not less than that of sound.
§89 Weak discontinuities 345
Weak discontinuities differ fundamentally from strong ones in the manner
of their occurrence. We shall see that shock waves can be formed as a direct
result of the gas flow, the boundary conditions being continuous (for instance,
the formation of shock waves in a sound wave, §95). In contrast to this, weak
discontinuities cannot occur spontaneously; they are always the result of
some singularity of the initial or boundary conditions of the flow. These
singularities may be of various kinds, like the weak discontinuities themselves.
For example, a weak discontinuity may occur on account of the presence of
angles on the surface of a body past which the flow takes place; in this case
the first spatial derivatives of the velocity are discontinuous. A weak dis
continuity is also formed when the curvature of the surface of the body is
discontinuous, without there being an angle; in this case the second spatial
derivatives of the velocity are discontinuous, and so on. Finally, any sin
gularity in the time variation of the flow results in a nonsteady weak dis
continuity.
The gas velocity component tangential to the surface of a weak discon
tinuity is always directed away from the point (e.g. an angle on the surface
of a body) from which the perturbation begins which causes the discontinuity;
we shall say that the discontinuity begins from this point. This is an example
of the fact that, in a supersonic flow, perturbations are propagated down
stream.
The presence of viscosity and thermal conduction results in a finite
thickness of a weak discontinuity, which is therefore in reality a transition
layer, like a shock wave. The thickness of the latter, however, depends only
on its strength and is constant in time, whereas the thickness of a weak
discontinuity increases with time after its formation. It is easy to determine
the law governing this increase. To do so, we again use the remark made at
the beginning of this section, that the motion of any part of the surface of a
weak discontinuity follows the same equations as the propagation of any weak
perturbation in the gas. When viscosity and thermal conduction occur, a
perturbation which is initially concentrated in a small volume (a "wave
packet") expands as it moves in the course of time; the manner of this expan
sion has been determined in §77. We can therefore conclude that the thick
ness S of a weak discontinuity is of the order of
8 ~ V(act), (89.1)
where t is the time from the formation of the discontinuity and a the co
efficient of the squared frequency in the sound absorption coefficient. If
the discontinuity is at rest, then the time t must be replaced by Ijc, where
/ is the distance from the point where the discontinuity starts (e.g. for a weak
discontinuity starting from an angle on the surface of a body, / is the distance
from the vertex of the angle); consequently 8 ~ \/(al). Thus the thickness
of a weak discontinuity increases as the square root of the time from its
formation or of the distance from its startingpoint.
To conclude this section, we should make the following remark, analogous
346 Shock Waves §89
to the one at the end of §79. We stated there that, among the various per
turbations of the state of a gas in motion, perturbations of entropy (at con
stant pressure) and vorticity are distinct in their properties. Such pertur
bations do not move relative to the gas, and are not propagated with the velocity
of sound. Hence the surfaces at which the entropy and vorticityf are weakly
discontinuous are at rest relative to the gas, and move with it relative to a
fixed system of coordinates. Such discontinuities may be called weak
tangential discontinuities; they pass through streamlines, and are in this respect
entirely analogous to the strong tangential discontinuities.
t A weak discontinuity of the vorticity implies a weak discontinuity of the velocity component
tangential to the surface of discontinuity; for example, the normal derivatives of the velocity may
be discontinuous.
CHAPTER X
ONEDIMENSIONAL GAS FLOW
§90. Flow of gas through a nozzle
Let us consider steady flow of a gas out of a large vessel through a tube of
variable crosssection (a nozzle). We shall suppose that the gas flow is
uniform over the crosssection at every point in the tube, and that the velocity
is along the axis of the tube. For this to be so, the tube must not be too wide,
and its crosssectional area S must vary fairly slowly along its length. Thus
all quantities characterising the flow will be functions only of the coordinate
along the axis of the tube. Under these conditions we can apply the relations
obtained in §80, which are valid along streamlines, directly to the variation
of quantities along the axis.
The mass of gas passing through a crosssection of the tube in unit time
(the discharge) is Q = pvS\ this must evidently be constant along the tube:
Q = Spv = constant. (90.1)
The linear dimensions of the vessel are supposed very large in comparison
with the diameter of the tube. The velocity of the gas in the vessel may there
fore be taken as zero, and accordingly all quantities with the suffix in the
formulae of §80 will be the values of those quantities in the vessel.
We have seen that the flux density j = pv cannot exceed a certain limiting
value j\. It is therefore clear that the possible values of the total discharge Q
have (for a given tube and a given state of the gas in the vessel) an upper
limit £ max , which is easily determined. If the value j m of the flux density
were reached anywhere except at the narrowest point of the tube, we should
have j > /„. for crosssections with smaller S, which is impossible. The
valued = jx can therefore be attained only at the narrowest point of the tube;
let the crosssectional area there be S min . Then the upper limit to the total
discharge is
£max = p.©.Snin = V(^0/>o)[2/(y+ l)]a+r)/2(yl) /S ' mln . (QQ.2)
Let us first consider a nozzle which narrows continually towards its outer
end, so that the minimum crosssectional area is at that end (Fig. 54). By
(90.1), the flux density j increases monotonically along the tube. The same is
true of the gas velocity v, and the pressure accordingly falls monotonically.
The greatest possible value of j is reached if v attains the value c just at the
outer end of the tube, i.e. if v\ — c\ = v* (the suffix 1 denotes quantities
pertaining to the outer end). At the same time, p\= p*.
Let us now follow the change in the manner of outflow of the gas when the
external pressure p e diminishes. When this pressure decreases from po, the
347
348
Onedimensional Gas Flow
§90
pressure inside the vessel, to />*, the pressure pi at the outer end of the tube
decreases also, and the two pressures pi andp e remain equal; that is, the whole
of the pressure drop from p to p e occurs in the nozzle. The velocity vi
with which the gas leaves the tube, and the total discharge Q = j\S min ,
increase monotonically, however. For p e = p* the velocity becomes equal to
the local velocity of sound, and the discharge reaches the value Q m&x . When
the external pressure decreases further, the pressure pi remains constant
at p*, and the fall of pressure from p* to p e occurs outside the tube, in the
surrounding medium. In other words, the pressure drop along the tube
cannot be greater than from_p to/)*, whatever the external pressure. For air
(p* = 053p ), the maximum pressure drop is 047/> . The velocity at the
end of the tube and the discharge also remain constant for p e < p*. Thus the
gas cannot acquire a supersonic velocity in flowing through a nozzle of this
kind.
Fig. 54
'£&//////<
Fig. 55
If we consider only the flow in the immediate neighbourhood of the end of
the tube, the motion of the gas after leaving the tube is essentially flow round
an angle, the vertex of which is the edge of the tube mouth; we shall discuss
this flow in detail in §104.
The impossibility of achieving supersonic velocities by flow through a
continually narrowing nozzle is due to the fact that a velocity equal to the
local velocity of sound can be reached only at the very end of such a tube. It
is clear that a supersonic velocity can be attained by means of a nozzle which
first narrows and then widens again (Fig. 55). This is called a de Laval
nozzle.
§90
Flow of gas through a nozzle
349
The maximum flux density j%, if reached, can again occur only at the
narrowest crosssection, so that the discharge cannot exceed S mi3 J^. In the
narrowing part of the nozzle, the flux density increases (and the pressure
falls); the curve in Fig. 56 shows j as a functionf of^>, and the variation just
described corresponds to the interval from c to b. If the maximum flux
density is reached at the crosssection S mln (the point b in Fig. 56), the pres
sure continues to diminish in the widening part of the nozzle, while^' begins to
decrease also, corresponding to the segment ba of the curve. At the outer
end of the tube; takes a definite value, / lmax = /* S ml JS h and the pressure
has the corresponding value, denoted in Fig. 56 by^i', at some point d on the
curve. If, however, only some point e is reached at the crosssection S mln ,
the pressure increases in the widening part of the nozzle, corresponding to a
return down the curve from e towards c. At first sight it might appear that we
might pass discontinuously from cb to ab, without going through the point b,
by the formation of a shock wave. This, however, is impossible, since the
gas "entering" the shock wave cannot have a subsonic velocity.
Bearing in mind these results, let us now investigate the manner of variation
in the outflow when the external pressure p e is gradually increased. For small
pressures, from zero to pi, the pressure p* and velocity v* = c* are reached
at the crosssection ,S min . In the widening part of the nozzle the velocity
continues to increase, so that there results a supersonic flow of the gas, and
the pressure accordingly continues decreasing, reaching the value pi at
the outer end of the tube, whatever the pressure p e . The pressure falls from
pi top e outside the nozzle, in the rarefaction wave which leaves the edge of
the tube mouth (see §104).
When p e exceeds pi, an oblique shock wave leaves the edge of the tube
mouth, compressing the gas from pi to p e (§104). We shall see, however,
that a steady shock wave can leave a solid surface only if its intensity is not too
f According to formulae (80.1580.17), the dependence is
(s)
2y
y1
popo
[(
.(yl)/y
po) II '
11*.
350 Onedimensional Gas Flow §91
great (§103). Hence, when the external pressure increases further, the shock
wave soon begins to move into the nozzle, with separation occurring in front
of it on the inner surface of the tube. For some value of p e the shock wave
reaches the narrowest crosssection and then disappears; the flow becomes
everywhere subsonic, with separation on the walls of the widening part of
the nozzle. All these complex phenomena are, of course, threedimensional.
PROBLEM
A small amount of heat is supplied over a short segment of a tube to a perfect gas in steady
flow in the tube. Determine the change in the gas velocity when it passes through this seg
ment.
Solution. Let Sq be the amount of heat supplied per unit time, S being the crosssectional
area of the tube at the segment concerned. The mass flux density j = pv and the momentum
flux density p+jv are the same on both sides of the heated segment; hence A/> = — jAy,
where A denotes the change in a quantity in passing through the segment. The difference in
the energy flux density (w+fr^j is q. Writing to = y/>/(yl)p = ypvl(y\)j, we obtain
(supposing Av and Ap small) vjAv + y(pAv+vAp)f(yl) = q. Eliminating Ap, we find
A„ _ (y—l)q/p(c 2 —v 2 ). We see that, in subsonic flow, the supply of heat accelerates the
flow (Av > 0), while in supersonic flow it retards it.
Writing the gas temperature asT= ftpJRp = ppv/Rj (R being the gas constant), we find
For supersonic flow, this expression is always positive, and the gas temperature is increased ;
for subsonic flow, however, AT may be either positive or negative.
§91. Flow of a viscous gas in a pipe
Let us consider the flow of a gas in a pipe (of constant crosssection) so
long that the friction of the gas against the walls, i.e. the viscosity of the gas,
cannot be neglected. We shall suppose the walls to be thermally insulated, so
that there is no heat exchange between the gas and the surrounding medium.
For gas velocities of the order of or exceeding the velocity of sound (the
only case we shall discuss here), the gas flow in the pipe is, of course, turbulent
if the radius of the pipe is not small. The turbulence of the flow is important,
as regards our problem, only in one respect: we have seen in §43 that, in
turbulent flow, the (mean) velocity is practically the same almost everywhere
in the crosssection of the pipe, and falls rapidly to zero very close to the
walls. We shall therefore suppose that the gas velocity v is a constant over
the crosssection, and define it so that the product Spv (S being the cross
sectional area) is equal to the total discharge through the crosssection.
Since the total discharge Spv is constant along the pipe, and S is assumed
constant, the mass flux density must also be constant:
j = pv — constant. (91.1)
Next, since the pipe is thermally insulated, the total energy flux carried by
the gas through any crosssection must also be constant. This flux is
Spv(w + ^v 2 '), and by (91.1) we have
w +%v 2 = w+^V 2 = constant. (91.2)
§91 Flow of viscous gas in a pipe 351
The entropy s of the gas does not, of course, remain constant, but increases
as the gas moves along the pipe, because of the internal friction. If x is
the coordinate along the pipe, with x increasing downstream, we can write
dsjdx > 0. (91.3)
We now differentiate (91.2) with respect to x. Since dzv = Tds+ Vdp, we
have
ds dp dV
T T +V T + J 2V J = 
dx dx dx
Next, substituting
dV
dx
we obtain
/ dV\ dp / 8V\ ds
= \~di) s dx + \17) p dx' (91 * 4)
h'^Js^h^JS (9L5)
By a wellknown formula of thermodynamics, (dV/ds) p = (TJc p )(dVI8T) p .
The coefficient of thermal expansion is positive for gases. We therefore
conclude, using (91.3), that the lefthand side of (91.5) is positive. The sign
of the derivative dpjdx is therefore that of [l+j 2 (dVldp) s ] = (vjcf\.
We see that
dp/d*§0 for v$c. (91.6)
Thus, in subsonic flow, the pressure decreases downstream, as for an in
compressible fluid. For supersonic flow, however, it increases.
We can similarly determine the sign of the derivative dvjdx. Since
j = vjV = constant, the sign of da/cbc is the same as that of dVJdx. The
latter can be expressed in terms of the positive derivative dsjdx by means of
(91.4) and (91.5). The result is that
dvjdx^O for v$c, (91.7)
i.e. the velocity increases downstream for subsonic flow and decreases for
supersonic flow.
Any two thermodynamic quantities for a gas flowing in a pipe are functions
of one another, independent of (inter alia) the resistance law for the pipe.
These functions depend on the constant j as a parameter, and are given by
the equation w + %j 2 V 2 = constant, which is obtained by eliminating the velo
city from the equations of conservation of mass and energy for the gas.
Let us ascertain the nature of the curves giving, for example, the entropy
as a function of pressure. Rewriting (91.5) in the form
ds (vjc) 2 — 1
dp T+jW(dVfds) p '
we see that, at the point where v = c, the entropy has an extremum. It is easy
352
Onedimensional Gas Flow
§91
to see that s has a maximum. For the second derivative of s with respect to
p at this point is
rd%
Ldp2
J V=C
]W{?W\dp*) s
<0;
T+pV(dVl8s) p
we assume, as usual, that the derivative {d 2 Vjdp 2 ) s is positive.
The curves giving $ as a function of p (called Fanno lines) are therefore as
shown in Fig. 57. The region of subsonic velocities lies to the right of the
maximum, and that of subsonic velocities to the left. When the parameter j
increases, we go to lower curves. For, differentiating equation (91.2) with
respect to j for constant p, we have
ds _ jV*
d/ = " T+jW(dV/ds) p
< 0.
Fig. 57
We can draw an interesting conclusion from the above results. Let the gas
velocity at the entrance to the pipe be less than that of sound. The entropy
increases downstream, and the pressure decreases; this corresponds to a
movement along the righthand branch of the curve s = s(p), from B to
wards O (Fig. 57). This can, however, continue only until the entropy reaches
its maximum value. A further movement along the curve beyond O (i.e. into
the region of supersonic velocities) is not possible, since the entropy of the
gas would have to decrease as it moved along the pipe. The transition be
tween the branches BO and OA cannot even be effected by a shock wave,
since the gas entering a shock wave cannot move with subsonic velocity.
Thus we conclude that, if the gas velocity at the entrance to the pipe is less
than that of sound, the flow remains subsonic everywhere in the pipe. The
gas velocity becomes equal to the local velocity of sound only at the other end
of the pipe, if at all (it does so if the pressure of the external medium into
which the gas issues is sufficiently low).
§92 Onedimensional similarity flow 353
In order that the gas should have supersonic velocities in the pipe, its
velocity at the entrance must be supersonic. By the general properties of
supersonic flow (the impossibility of propagating disturbances upstream),
the flow will then be entirely independent of the conditions at the outlet of
the pipe. In particular, the entropy will increase along the pipe in a quite
definite manner, and its maximum value will be attained at a definite distance
x = h from the entrance. If the total length / of the pipe is less than 4, the
flow is supersonic throughout the pipe (corresponding to movement on the
branch AO from A towards O). If, on the other hand, / > l k , the flow cannot
be supersonic throughout the pipe, nor can there be a smooth transition to
subsonic flow, since we can move along the branch OB only in the direction
shown by the arrow. In this case, therefore, a shock wave must necessarily
be formed, which discontinuously changes the flow from supersonic to sub
sonic. The pressure is thereby increased, and we pass from the branch AO
to BO without going through the point O. The flow is entirely subsonic
beyond the discontinuity.
§92. Onedimensional similarity flow
An important class of onedimensional nonsteady gas flows is formed by
flows occurring in conditions where there are characteristic velocities but
not characteristic lengths. The simplest example of such a flow is given by
gas flow in a semiinfinite cylindrical pipe terminated by a piston, when the
piston begins to move with constant velocity.
Such a flow is defined by the velocity parameter and by parameters which
give, say, the gas pressure and density at the initial instant. We can, however,
form no combination of these parameters which has the dimensions of length
or time. It therefore follows that the distributions of all quantities can depend
on the coordinate x and the time t only through the ratio xft, which has the
dimensions of velocity. In other words, these distributions at various in
stants will be similar, differing only in the scale along the xaxis, which in
creases proportionally to the time. We can say that, if lengths are measured
in a unit which increases proportionally to t, then the flow pattern does not
change. When the flow pattern is unchanged with time if the scale of length
varies appropriately, we speak of a similarity flow.
The equation of conservation of entropy for a flow which depends on only
one coordinate, *, is dsjdt + Vx ds/dx = 0. Assuming that all quantities
depend only on £ = xjt, and noticing that in this case d/dx = (l/f)d/df,
djdt = (£/*)d/d£, we obtain (v x  £) s' = (the prime denoting differen
tiation with respect to £). Hence s' = 0, i.e. s = constantf ; thus similarity
flow in one dimension must be isentropic. Likewise, from the y and z com
ponents of Euler's equation: dv y ldt + v x dv y jdx = 0, dv z /dt + v x dv z /dx = 0,
we find that v y and v z are constants, which we can take as zero without loss
of generality.
t The assumption that v x — £ = would contradict the other equations of motion; from (92.3)
we should have v x = constant, contrary to hypothesis.
354 Onedimensional Gas Flow §92
Next, the equation of continuity and the ^component of Euler's equation
are
dp dv dp
■£ + T + V = ' (92<1)
dt ox ox
dv dv 1 dp
+ V= __Z ; ( 92.2)
dt ox p ox
here and henceforward we write v x as v simply. In terms of the variable $ ,
these equations become
(v€)p'+pv' = 0, (92.3)
( V Qv> = p'j P = cy/p. (92.4)
In the second equation we have putp' = {dp\dp) s p = c 2 p', since the entropy
is constant.
These equations have, first of all, the trivial solution v = constant,
p = constant, i.e. a uniform flow of constant velocity. To find a nontrivial
solution, we eliminate p and v' from the equations, obtaining (v— £) 2 = c 2 y
whence  = v± c. We shall take the plus sign:
x\t = v + c; (92.5)
this choice of sign means that we take the positive #axis in a definite direction,
selected in a manner shown later. Finally, putting v— £ = — c in (92.3),
we obtain cp = pv', or pdv = cdp. The velocity of sound is a function of
the thermodynamic state of the gas; taking as the fundamental thermodyna
mic quantities the entropy s and the density />, we can represent the velocity
of sound as a function c(p) of the density, for any given value of the constant
entropy. With c understood as such a function, we can write
v = j cdp/p = j dpjcp. (92.6)
This formula can also be written
v = j ^/(dpdV), (92.7)
in which the choice of dependent variable remains open.
Formulae (92.5) and (92.6) give the required solution of the equations of
motion. If the function c(p) is known, then the velocity v can be calculated
as a function of density from (92.6). Equation (92.5) then determines the
density as an implicit function of xjt, and so the dependence of all the other
quantities on xjt is determined also.
We can derive some general properties of the solution thus obtained.
Differentiating equation (92.5) with respect to x, we have
dp d(v + c)
t— \ = 1. (92.8)
dx dp
§92 Onedimensional similarity flow 355
For the derivative of v + c we have, by (92.6),
d(v + c) c dc 1 d(pc)
dp p dp p dp
But
P c = pvWp) = W(W;
differentiating, we have
dOoc)/dp = &l(pc)ldp = lp*(*(d*VldpP) 8 . (92.9)
Thus
d(© + c)/dp = y 2 c%8W/dp^) s > 0. (92.10)
It therefore follows from (92.8) that dp/dx > for t > O.f Since dp/dx
= c 2 dp/dx, we conclude that dp/&* > also. Finally, we have dvjdx
= (c/p)dp/dx, so that dvjdx > 0. The inequalities
8p/8x > 0, dp/dx > 0, dv/dx > (92.11)
therefore hold.
The meaning of these inequalities becomes clearer if we follow the variation
of quantities, not along the xaxis for given t, but with time for a given gas
element as it moves about. This variation is given by the total time deriva
tive ; for the density, for example, we have, using the equation of continuity,
dp/dt = dp/dt+v dp/dx = —p dv/dx. By the third inequality (92.11),
this quantity is negative, and therefore so is dp/dt:
dp/dt < 0, dp/dt < 0. (92.12)
Similarly (using Euler's equation (92.2)) we can see that dv/dt < 0; this,
however, does not mean that the magnitude of the velocity diminishes with
time, since v may be negative.
The inequalities (92.12) show that the density and pressure of any gas
element decrease as it moves. In other words, the gas is continually rarefied
as it moves. Such a flow may therefore be called a nonsteady rarefaction
wave.
A rarefaction wave can be propagated only a finite distance along the #axis ;
this is seen from the fact that formula (92.5) would give an infinite velocity
for x > + oo, which is impossible.
Let us apply formula (92.5) to a plane bounding the region of space occupied
by the rarefaction wave. Here x/t is the velocity of this boundary relative to
the fixed coordinate system chosen. Its velocity relative to the gas itself is
(x/t) — v and is, by (92.5), equal to the local velocity of sound. This means
that the boundaries of a rarefaction wave are weak discontinuities. The
t There is no meaning for times t < in the similarity flow here considered. Such a flow can
occur only because of some singularity in the initial conditions (t = 0) of the flow at the point x = 0,
and therefore takes place only for t > (in our example, the piston velocity changes discontinuously
at t  0. See also §93).
356 Onedimensional Gas Flow §92
similarity flow in different cases is therefore made up of rarefaction waves and
regions of constant flow, separated by surfaces of weak discontinuity.  ]
The choice of sign in (92.5) is now seen to correspond to the fact that these
weak discontinuities are assumed to move in the positive ^direction relative
to the gas. The inequalities (92.11) arise from this choice, but the inequalities
(92.12), of course, do not depend on the direction of the #axis.
r
J
m !
I
I
j
IT
Fig. 58
We are usually concerned, in actual problems, with a rarefaction wave
bounded on one side by a region where the gas is at rest. Let this region (I
in Fig. 58) be to the right of the rarefaction wave. Region II is the rarefaction
wave, and region III contains gas moving with constant velocity. The arrows
in the figure show the direction of motion of the gas, and of the weak dis
continuities bounding the rarefaction wave; the discontinuity a always
moves into the gas at rest, but the discontinuity b may move in either direction,
depending on the velocity reached in the rarefaction wave (see Problem 2).
We may give explicitly the relations between the various quantities in such a
rarefaction wave, assuming that we have a perfect gas. For an adiabatic
process pT ll(1 ~ y) = constant. Since the velocity of sound is proportional to
\/T, we can write this relation as
P = po{chf'^\ (92.13)
Substituting this expression in the integral (92.6), we obtain
2 f 2
v = dc = (cco);
y—l J y— 1
the constant of integration is chosen so that c = Co for v = (we use the
suffix to refer to the point where the gas is at rest). We shall express all
quantities in terms of v, bearing in mind that, with the above situation of the
various regions, the gas velocity is in the negative ^direction, i.e. v < 0.
Thus
c = c Q \{y\)\v\, (92.14)
which determines the local velocity of sound in terms of the gas velocity.
Substituting in (92.13), we find the density to be
P = Po[l¥yl)\v\lco]Wy», (92.15)
f There may also, of course, be regions of constant flow separated by shock waves.
§92
Onedimensional similarity flow 357
and similarly the pressure is
P = Mi \{y~ i)M M)] 2 ^ 1 *. (92.16)
Finally, substituting (92.14) in formula (92.5), we obtain
2 / x
M = 7 Ko  
1 ' y+l\ t
(92.17)
which gives a as a function of # and t.
The quantity c cannot be negative, by definition. We can therefore draw
from (92.14) the important conclusion that the velocity must satisfy the
inequality
M <2* /(yl); ( 92  18 )
when the velocity reaches this limiting value, the gas density (and also/) and
c) becomes zero. Thus a gas originally at rest and expanding nonsteadily
in a rarefaction wave can be accelerated only to velocities not exceeding
2*o/(yl). . . r L . • • ,
We have already mentioned, at the beginning of this section, a simple
example of similarity flow, namely that which occurs in a cylindrical pipe in
which a piston begins to move with constant velocity. If the piston moves out
of the pipe, it creates a rarefaction, and a rarefaction wave of the kind des
cribed above is formed. If, however, the piston moves inwards, it compresses
the gas in front of it, and the transition to the original lower pressure can occur
only in a shock wave, which is in fact formed in front of a piston moving for
ward in a pipe (see the following Problems)!
PROBLEMS
Problem 1. A perfect gas occupies a semiinfinite cylindrical pipe terminated by a
piston. At an initial instant the piston begins to move into the pipe with constant velocity U.
Determine the resulting flow.
Solution A shock wave is formed in front of the piston, and moves along the pipe.
At the initial instant this shock and the piston are coincident, but at subsequent instants the
shock is ahead of the piston, and a region of gas lies between them (region 2). In front or
the shock wave (region 1), the gas pressure is equal to its initial value p u and its velocity
relative to the pipe is zero. In region 2, the gas moves with constant velocity, equal to the
velocity U of the piston (Fig. 59). The difference in velocity between regions 1 and 2 is
therefore also U, and, by formulae (82.7) and (85.1), we can write
u= V[(p2piWiv 2 )]
= (p2piW{2Vil[(yl)Pi + (y+^)p2]y
t We may mention also an analogous similarity flow in three dimensions: the centrally symmetrical
sas flow caused by a uniformly expanding sphere. A spherical shock wave, expanding with constant
velocity is formed in front of the sphere. Unlike what happens in the onedimensional case, the
velocity of the gas between the sphere and the shock is not constant; the equation which determines
it as a function of the ratio r\t (and therefore the rate of propagation of the shock wave) cannot be
m Twf ptoWemha^been discussed by L. I. SedoV (1945; see his book Similarity and Dimensional
Methods in Mechanics, CleaverHume Press, London 1959) and by G. I. Taylor, Proceedings of the
Royal Society, A186, 273, 1946.
358
Onedimensional Gas Flow
§92
Hence we find the gas pressure p z between the piston and the shock wave to be given by
( y+ l)2f/2
— X T " 1 / 1 t
Pi
4c,2
Cl
\6a*
Knowing p 2 , we can calculate, from formulae (85.4), the velocity of the shock wave relative
to the gas on each side of it. Since gas 1 is at rest, the velocity of the shock relative to it is
equal to the rate of propagation of the shock in the pipe. If the x coordinate (along the pipe)
is measured from the initial position of the piston (the gas being on the side x > 0), we find
the position of the shock wave at time t to be
while the position of the piston is x = Ut.
t
U
Fig. 59
(a) ^
2 3
U
(fo^0) f
Problem 2. The same as Problem 1, but for the case where the piston moves out of the
pipe with velocity U.
Solution. The piston adjoins a region of gas (region 1 in Fig. 60a) which moves in the
negative ^direction with constant velocity — U, equal to the velocity of the piston. Then
follows a rarefaction wave (2), in which the gas moves in the negative ^direction, its velocity
varying linearly from — U to zero according to (92.17). The pressure varies according to
(92.16) from p x = £ [l_£( y _l)£// Co ]2y/ ( yi) in gas j to Po in the gas 3> which fg &t res ^
The boundary of regions 1 and 2 is given by the condition v — —U; according to (92.17),
we have x = [c —}(y+l)U]t = (c—U)t, where c is the velocity of sound in gas 1. At the
boundary of regions 2 and 3, v = 0, whence * = c t. Both boundaries are weak discon
tinuities ; the second is always propagated to the right (i.e. away from the piston), but the first
may be propagated either to the right (as shown in Fig. 60a) or to the left (if the piston
velocity U > 2c /(y+l)).
§92
Onedimensional similarity flow
359
The flow pattern just described can occur only if U < 2c /(y — 1). If U > 2c„/(y — 1), a
vacuum is formed in front of the piston (the gas cannot follow the piston), which extends
from the piston to the point x = 2c «/(yl) (region 1 in Fig. 60b). At this point,
v = 2c /(yl); then follow region 2, in which the velocity decreases to zero at the point
x = c t, and region 3, where the gas is at rest.
Problem 3. A gas occupies a semiinfinite cylindrical pipe (x > 0) terminated by a valve.
At time t = 0, the valve is opened, and the gas flows into the external medium, the pressure
pe in which is less than the initial pressure p in the pipe. Determine the resulting flow.
Fig. 61
Solution. Let — v e be the gas velocity which corresponds to the external pressure p e
according to formula (92.16); for * = and t > 0, we must have v = v e . If
v e < 2c /(y+l), the velocity distribution shown in Fig. 61a results. For v e = 2c /(y+l)
(corresponding to a rate of outflow equal to the local velocity of sound at the end of the pipe :
this is easily seen by putting v = c in formula (92.14)), the region of constant velocity vanishes
and the pattern shown in Fig. 61b is obtained. The quantity 2c /(y + l) is the greatest
possible rate of outflow from the pipe in the conditions stated. If the external pressure pe
is such that __„ . . /1N
p e </>o[2/(y + l)] 2 >' /( ^ 1) , (1)
the corresponding velocity exceeds 2c /(y+l). In reality, the pressure at the pipe outlet
would still be equal to the limiting value (the righthand side of (1)), and the rate of outflow
would be 2c„/(y+l); the remaining pressure drop (to p e ) occurs in the external medium.
Problem 4. An infinite pipe is divided by a piston, on one side of which (x < 0) there is,
at the initial instant, gas at pressure p , and on the other side a vacuum. Determine the motion
of the piston as the gas expands.
Solution. A rarefaction wave is formed in the gas ; one of its boundaries moves to the right
with the piston, and the other moves to the left. The equation of motion of the piston is
mdU/dt = po[\\{y\)Ujc Q fy^^\
where U is the velocity of the piston and m its mass per unit area. Integrating, we obtain
™SH'^]
(y+l)po\^ my+1) \
Problem 5. Determine the flow in an isothermal similarity rarefaction wave.
Solution. The isothermal velocity of sound is ct = V(dp/dp)T = V(RTffi), and for
constant temperature ct = constant = cto According to (92.5) and (92.6), we therefore
v = c To log(p/po) = c T ^og{pjpo) = (xji)c Tt .
360 Onedimensional Gas Flow §93
§93. Discontinuities in the initial conditions
One of the most important reasons for the occurrence of surfaces of dis
continuity in a gas is the possibility of discontinuities in the initial conditions.
These conditions (i.e. the initial distributions of velocity, pressure, etc.)
may in general be prescribed arbitrarily. In particular, they need not be
everywhere continuous, but may be discontinuous on various surfaces. For
example, if two masses of gas at different pressures are brought together at
some instant, their surface of contact will be a surface of discontinuity of the
initial pressure distribution.
It is of importance that the discontinuities of the various quantities in the
initial conditions (or, as we shall say, in the initial discontinuities) can have
any values whatever; no relation between them need exist. We know, how
ever, that certain conditions must hold on stable surfaces of discontinuity in a
gas; for instance, the discontinuities of density and pressure in a shock wave
are related by the shock adiabatic. It is therefore clear that, if these conditions
are not satisfied in the initial discontinuity, it cannot continue to be a dis
continuity at subsequent instants. Instead, the initial discontinuity in general
splits into several discontinuities, each of which is one of the possible types
(shock wave, tangential discontinuity, weak discontinuity) ; in the course of
time, these discontinuities move apart. A general discussion of the behaviour
of an arbitrary discontinuity has been given by N. E. Kochin (1926).
During a short interval of time after the initial instant t = 0, the discon
tinuities formed from the initial discontinuity do not move apart to great
distances, and the flow under consideration therefore takes place in a relatively
small volume adjoining the surface of initial discontinuity. As usual, it
suffices to consider separate portions of this surface, each of which may be
regarded as plane. We need therefore consider only a plane surface of
discontinuity, which we take as the yzplzne. It is evident from symmetry
that the discontinuities formed from the initial discontinuity will also be plane,
and perpendicular to the *axis. The flow pattern will depend on the co
ordinate x only (and on the time), so that the problem is onedimensional.
There being no characteristic parameters of length and time, we have a
similarity problem, and the results obtained in §92 can be used.
The discontinuities formed from the initial discontinuity must evidently
move away from their point of formation, i.e. away from the position of the
initial discontinuity. It is easy to see that either one shock wave, or one pair
of weak discontinuities bounding a rarefaction wave, can move in each direc
tion (the positive and negative ^direction). For, if there were, say, two shock
waves formed at the same point at time t = and both propagated in the
positive xdirection, the leading one would have to move more rapidly than
the other. According to the general properties of shock waves, however, the
leading shock wave must move, relative to the gas behind it, with a velocity
less than the velocity of sound c in that gas, and the following shock must
move, relative to the same gas, with a velocity exceeding c (c being a constant
in the region between the shock waves), i.e. it must overtake the other. For
§93 Discontinuities in the initial conditions 361
the same reason, a shock wave and a rarefaction wave cannot move in the same
direction; to see this, it is sufficient to notice that weak discontinuities move
with the velocity of sound relative to the gas on each side of them. Finally,
two rarefaction waves formed at the same time cannot become separated,
since the velocities of their backward fronts are the same.
As well as shock waves and rarefaction waves, a tangential discontinuity
must in general be formed from an initial discontinuity. Such a discontinuity
must occur if the transverse velocity components v y , v z are discontinuous in
the initial discontinuity. Since these velocity components do not change in a
shock or rarefaction wave, their discontinuities always occur at a tangential
discontinuity, which remains at the position of the initial discontinuity; on
each side of this discontinuity, v y and v z are constant (in reality, of course,
the instability of a tangential velocity discontinuity causes its gradual smooth
ing into a turbulent region).
A tangential discontinuity must occur, however, even if v y and v z are
continuous at the initial discontinuity (without loss of generality, we can, and
shall, assume that they are zero). This is shown as follows. The discon
tinuities formed from the initial discontinuity must make it possible to go from
a given state 1 of the gas on one side of the initial discontinuity to a given state
2 on the other side. The state of the gas is determined by three independent
quantities, e.g. p, p and v x = v. It is therefore necessary to have three arbi
trary parameters in order to go from state 1 to an arbitrary state 2 by some
choice of the discontinuities. We know, however, that a shock wave, per
pendicular to the stream, propagated in a gas whose thermodynamic state is
given, is completely determined by one parameter (§82). The same is true of
a rarefaction wave; as we see from formulae (92.14)(92.16), when the state
of the gas entering a rarefaction wave is given, the state of the gas leaving it is
completely determined by one parameter. We have seen, moreover, that at
most one wave (rarefaction or shock) can move in each direction. We therefore
have at our disposal only two parameters, which are not sufficient.
The tangential discontinuity formed at the position of the initial discon
tinuity furnishes the third parameter required. The pressure is continuous
there, but the density (and therefore the temperature and entropy) is not.
The tangential discontinuity is stationary with respect to the gas on both sides
of it and the arguments about the "overtaking" of two waves propagated in the
same direction therefore do not apply to it.
The gases on the two sides of the tangential discontinuity do not mix,
since there is no motion of gas through a tangential discontinuity; in all the
examples given below, these gases may be different substances.
Fig. 62 shows schematically all possible types of breakup of an initial
discontinuity. The continuous line shows the variation of the pressure along
the #axis; the variation of the density would be given by a similar line, the
only difference being that there would be a further jump at the tangential
discontinuity. The vertical lines show the discontinuities formed, and the
arrows show their direction of propagation and that of the gas flow. The
362
Onedimensional Gas Flow
§93
coordinate system is always that in which the tangential discontinuity is at
rest, together with the gas in the regions 3 and 3' which adjoin it. The pres
sures, densities and velocities of the gases in the extreme lefthand (1) and
righthand (2) regions are the values of these quantities at time t = on each
side of the initial discontinuity.
3 3
2
1
—
*
*~S+TS^ , %
(a)
3
3'
4/
1 •"•
**" !
—
I*S^_TR_
(b)
1
3'
r !

1 • '
— i — i i_
1 + R^.TR
(c)
ZW?
r "*■ (d)
Shock wave
Tangential discontinuity
■ Weak discontinuity
Fig. 62
In the first case, which we write I > S^ TS_* (Fig. 62a), the initial dis
continuity / gives two shock waves S, propagated in opposite directions,
and a tangential discontinuity T between them. This case occurs when two
masses of gas collide with a large relative velocity.
In the case / > fi_ TR_> (Fig. 62b), a shock wave is propagated on one
side of the tangential discontinuity, and a rarefaction wave R on the other
side. This case occurs, for instance, if two masses of gas at relative rest
(v2 — V! = 0) and at different pressures are brought into contact at the initial
§93 Discontinuities in the initial conditions 363
instant. For, of all the cases shown in Fig. 62, the second is the only one in
which gases 1 and 2 are moving in the same direction, and so the equation
vi — v 2 is possible. .
In the third case (/ > i?_ TR^, Fig. 62c), a rarefaction wave is propagated
on each side of the tangential discontinuity. If gases 1 and 2 separate with a
sufficiently great relative velocity v 2 v x , the pressure may decrease to zero
in the rarefaction waves. We then have the pattern shown in Fig. 62d; a
vacuum 3 is formed between regions 4 and 4'.
We can derive the analytical conditions which determine the manner in
which the initial discontinuity breaks up, as a function of its parameters. We
shall suppose in every case that£ 2 > pi, and take the positive ^direction from
region 1 to region 2 (as in Fig. 62).
Since the gases on the two sides of the initial discontinuity may be ot
different substances, we shall distinguish them as gases 1 and 2.
(1) / * S<_TS^. If pz = pz', Vz and Vz' are the pressures and specific
volumes in the resulting regions 3 and 3', then we have pz>pz> pi, and the
volumes Vz and V Z ' are the abscissae of the points with ordinate pz on the
shock adiabatics through (p lf Vi) and (p 2 , V 2 ) respectively. Since the gases
in regions 3 and 3' are at rest in the coordinate system chosen, we can use
formula (82.7) to give the velocities v x and v 2 , which are in the positive and
negative ^directions respectively:
vi = VlipzpiWi v *)l v 2 =  V[(p8^2)(^2 Vz')].
The least value of pz, for given pi and p 2 , which does not contradict the initial
assumption (p 3 > p2 > pi) is pi Since, moreover, the difference ^  c* is a
monotonically increasing function of p z , we find the required inequality
viv 2 > Vmpi)(Vi V')], (93.1)
where V denotes the abscissa of the point with ordinate p 2 on the shock
adiabatic for gas 1 through (/>i, Vi). Calculating V from formula (85.1) (in
which V 2 is replaced by V), we obtain the condition (93.1) for a perfect gas
in the form
V!V2 > (P2/>i)V(2M(ni)/>i + (n + i)KI} ( 93  2 )
It should be noted that the limits placed by (93.1) and (93.2) on the possible
values of the velocity difference v 1 v 2 clearly do not depend on the coordi
nate system chosen.
(2) / > S^TR^. Here^i <pz =pz'<P* For the gas velocity m region 1
we again have
*>i = y/[{pzpiWiVz)l
and the total change in velocity in the rarefaction wave 4 is, by (92.7),
Pi
v 2 = J V(#dF).
P*
364 Onedimensional Gas Flow §93
For given p x and p 2 , p 3 can lie between them. Replacing p z in the difference
vzvx by pi and then by p 2 , we obtain the condition
 j V(dpdV) < V!v 2 < V[(pzpi){ViV')].
(93.3)
Here V has the same significance as in the previous case; the upper limit of
the difference V!  v 2 must be calculated for gas 1, and the lower limit for gas 2.
For a perfect gas we have
2c 2 r /p 1 \(y.i)/2y,i
< (P2 pi)V{2Vi/[(n  l)^i + (yi + l)pz]}, (93.4)
where c 2 = ^/(y 2 p 2 V 2 ) is the velocity of sound in gas 2 in the state (p 2 , V 2 ).
(3) / > R^TR^. Here p 2 > p 1 >p 3 = p z > > 0. By the same method we
find the following condition for this case to occur:
 j V(dpdV) J V(dpdV) < v x v 2 <  j ^/(dpdV). (93.5)
The first integral in the first member is calculated for gas 1, and the others
for gas 2. For a perfect gas we find
2ci 2c 2 2c 2 r / p x \ (y s i)/2y 2 i
7 <V!v 2 < h_(£_ I (93. 6 )
yi1 y 2 \ 721 L \p 2 ] y K '
where c ± = Vinpi^i), c 2 = V(72p2V 2 ). If
2c± 2c 2
viv 2 < , (93.7)
yi1 y2l
a vacuum is formed between the rarefaction waves (/ > R+_ VRJ).
The problem of a discontinuity in the initial conditions includes that
of various collisions between plane surfaces of discontinuity. At the instant of
collision, the two planes coincide, and form some initial discontinuity, which
then leads to one of the patterns described above. The collision of two shock
waves, for instance, results in two other shock waves, which move away
from the tangential discontinuity remaining between them : S^S*. » S+. TS^.
When one shock wave overtakes another, there are two possibilities: S_*S^
> S±_ TS_> and S_>S^ > R^ TS^. In either case a shock wave continues in the
same direction.
The problem of the reflection and transmission of shock waves by a tan
gential discontinuity (boundary of two media) also comes under this heading.
Here two cases are possible: S^T+S<.TS_ and S^T>R+_TS_+. The wave
§93
Discontinuities in the initial conditions
365
transmitted into the second medium is always a shock (see also the following
Problems)!
PROBLEMS
Problem 1. A plane shock wave is reflected from a rigid plane surface. Determine the gas
pressure behind the reflected wave (S. V. Izmailov 1935).
*■ 1
Fig. 63
/,
Solution When a shock wave is incident on a rigid wall, a reflected shock wave is pro
pagated away from the wall. We denote by the suffixes 1, 2 and 3 respectively quantities
pertaining to the undisturbed gas in front of the incident shock, the gas behind this shock
(which is also the gas in front of the reflected shock) and the gas behind the reflected shock;
see Fig. 63, where the arrows indicate the direction of motion of the shock waves and of the
gas itself. The gas in regions 1 and 3, which adjoin the wall, is at rest relative to the wall.
The relative velocity of the gases on the two sides of the discontinuity is the same in both
the incident and the reflected shock wave, and equal to the velocity of gas 2. Using formula
(82.7) for the relative velocity, we therefore have (p 2 — £i)(I / i — V 2 ) = (p 3 — p2)(.V 2 — V 3 ). The
equation of the shock adiabatic (85.1) for each shock gives
Yl
(y+l)/>i+(yl)/>2
Yl
v 2
(y+l)/>2+(yl)/*
(yl)/>i + (y+l)/>2 V 2 (yl)p2 + (y+l)^ 3
We can eliminate the specific volumes from these three equations, and the result is
(/>3/>2) 2 [(y+l)£i + (yl)/>2] = (p2£i) 2 [(y+l)/>3+(yl)M
This is a quadratic equation for p3, which has the trivial root p 3 = pi, cancelling p 3 — pu
we obtain
p3_ = (3yl)j>2(yl)j>i
P2 ~ (yl)/>2(y+l)^i'
which determines p 3 from p x and p t . In the limiting case of a very strong incident shock,
p 3 = (3yl)£ 2 /(yl), while for a weak shock p 3 p 2 = Pzpi, corresponding to the sound
wave approximation.
Problem 2. Find the condition for a shock wave to be reflected from a plane boundary
between two gases.
 For completeness we should mention that, when a shock wave collides with a weak discontinuity
(a problem which is not of the similarity type considered here), the shock wave continues to be
propagated in the same direction, but behind it there remain a weak discontinuity of the original
kind and a weak tangential discontinuity (see the end of §89).
366 OneDimensional Gas Flow §94
Solution. Let p^ = p r , V u V 2 ', be the pressures and specific volumes of the two media
before the incidence of the shock wave (propagated in gas 2), at their surface of separation,
and p 2 ,V 2 the values behind the shock wave. The condition for the reflected wave to be a
shock wave is given by the inequality (93.2), in which we must now put
Vl v 2 = V[(p2p2')(V 2 'V 2 )].
srms of the ratio of pressures p 2 /pi i
Expressing all quantities in terms of the ratio of pressures p % \p x and the initial specific volumes
V u V r , we obtain
(yi + l)p2lpi + (n 1) (Y2+l)p2lpi+(n 1)
§94. Onedimensional travelling waves
In discussing sound waves in §63, we assumed the amplitude of oscillations
in the wave to be small. The result was that the equations of motion were
linear and were easily solved. A particular solution of these equations is any
function of x±ct (a plane wave), corresponding to a travelling wave whose
profile moves with velocity c, its shape remaining unchanged; by the profile
of a wave we mean the distribution of density, velocity, etc., along the direc
tion of propagation. Since the velocity v, the density p and the pressure p
(and the other quantities) in such a wave are functions of the same quantity
x±ct y they can be expressed as functions of one another, in which the co
ordinates and time do not explicitly appear (p = p(p), v = v(p), and so on).
When the wave amplitude is not necessarily small, these simple relations
do not hold. It is found, however, that a general solution of the exact equa
tions of motion can be obtained, in the form of a travelling plane wave which is
a generalisation of the solution f(x ± ct) of the approximate equations valid
for small amplitudes. To derive this solution, we shall begin from the require
ment that, for a wave of any amplitude, the velocity can be expressed as a
function of the density.
In the absence of shock waves the flow is adiabatic. If the gas is homo
geneous at some initial instant (so that, in particular, s = constant), then
s = constant at all times, and we shall assume this in what follows.
In a plane sound wave propagated in the ^direction, all quantities depend
on x and t only, and for the velocity we have v x = v, v y = v z = 0. The
equation of continuity is 8p]8t+8(pv)[8x = 0, and Euler's equation is
8v 8v 1 8p
— + v— +  — = 0.
8t 8x p 8x
Using the fact that v is a function of p only, we can write these equations
as
8p d(pv) 8p
8v I 1 d*\ 8v
— + \v +f — = 0. (94.2)
8t \ pdv/8x K '
§94 Onedimensional travelling waves 367
Since
dpjdt (dx\
dp/dx ~ \dt)\
we have from (94.1)
(
and similarly from (94.2)
8x \ d(ov) dv
dt/ p dp dp
( d _l) =v +\% (94.3)
\8t/ v p dv
Since the value of p uniquely determines that of v, the derivatives for con
stant p and constant v are the same, i.e. (dx\dt) p = (dxjdt) v , so that p dvldp
= (1/p) dpldv. Putting dpjdv = (dj>/dp)(d/>/d«>) = ^dp/dv, we obtain dvjdp
= ± cjp, whence
r + f%±f£ (944)
This gives the general relation between the velocity and the density or pressure
in the wave.f
Next, we can combine (94.3) and (94.4) to give (Bxjdt) v = v + {\jp)dpldv
= v ± c(v), or, integrating,
x= t[v±c(v)]+f(v), (94.5)
where f{v) is an arbitrary function of the velocity, and c{v) is given by (94.4).
Formulae (94.4) and (94.5) give the required general solution (B. Riemann,
1860). They determine the velocity (and therefore all other quantities) as
an implicit function of x and t, i.e. the wave profile at every instant. For
any given value of v, we have * = at + b, i.e. the point where the velocity
has a given value moves with constant velocity; in this sense, the solution
obtained is a travelling wave. The two signs in (94.5) correspond to waves
propagated (relative to the gas) in the positive and negative ^directions.
The flow described by the solution (94.4) and (94.5) is often called a
simple wave, and we shall use this expression below. It should be noticed
that the similarity flow discussed in §92 is a particular case of a simple wave,
corresponding to f(v) = in (94.5).
We can write out explicitly the relations for a simple wave in a perfect
gas; for definiteness, we assume that there is a point in the wave for which
v = 0, as usually happens in practice. Since formula (94.4) is the same as
(92.6), we have by analogy with formulae (92.14)(92.16)
c = c ±l(yl)v, (94.6)
P = Po(i±Kri)^o) 2/<y  1) , (947)
/»=Mi±l(riW^o) 2 ^ 1) .
t In a wave of small amplitude we have p = Po +p, and (94.4) gives in the first approximation
v = Cop'lpo (where c = c(p )), i.e. the usual formula (63.12).
368 Onedimensional Gas Flow §49
Substituting (94.6) in (94.5), we obtain
x=t(± c +i(y+ l)v)+f(v). (94.8)
It is sometimes convenient to write this solution in the form
v = F[x(±c + i(y+l)v)t], (94.9)
where F is another arbitrary function.
From formulae (94.6) and (94.7) we again see (as in §92) that the velocity
in a direction opposite to that of the propagation of the wave (relative to the
gas itself) is of limited magnitude; for a wave propagated in the positive
^direction we have
v ^ 2c /(yl). (94.10)
A travelling wave described by formulae (94.4) and (94.5) is essentially
different from the one obtained in the limiting case of small amplitudes.
The velocity of a point in the wave profile is
u = v±c\ (94.11)
it may be conveniently regarded as a superposition of the propagation of a
disturbance relative to the gas with the velocity of sound and the movement
of the gas itself with velocity v. The velocity u is now a function of the
density, and therefore is different for different points in the profile. Thus,
in the general case of a plane wave of arbitrary amplitude, there is no definite
constant "wave velocity". Since the velocities of different points in the wave
profile are different, the profile changes its shape in the course of time.
Let us consider a wave propagated in the positive ^direction, for which
« = v + c. The derivative of v + c with respect to the density has been cal
culated in §92; see (92.10). We have seen that du/dp > 0. The velocity of
propagation of a given point in the wave profile is therefore the greater, the
greater the density. If we denote by c the velocity of sound for a density
equal to the equilibrium density p , then in compressions p > p and c> c ,
while in rarefactions p < p and c < c .
The inequality of the velocity of different points in the wave profile causes
its shape to change in the course of time: the points of compression move
forward and those of rarefaction are left behind (Fig. 64b). Finally, the
profile may become such that the function p{x) (for given t) is no longer
onevalued; three different values of p correspond to some x (the dashed
line in Fig. 64c). This is, of course, physically impossible. In reality,
discontinuities are formed where p is not one valued, and p is consequently
onevalued everywhere except at the discontinuities themselves. The wave
profile then has the form shown by the continuous line in Fig. 64c. The
surfaces of discontinuity are thus formed at points a wavelength apart.
When the discontinuities are formed, the wave ceases to be a simple wave.
The cause of this can be briefly stated thus: when surfaces of discontinuity
are present, the wave is "reflected" from them, and therefore ceases to be a
§94
Onedimensional travelling waves
369
wave travelling in one direction. The assumption on which the whole
derivation is based, namely that there is a onetoone relation between the
various quantities, consequently ceases to be valid in general.
The presence of discontinuities (shock waves) results, as was mentioned in
§82, in the dissipation of energy. The formation of discontinuities therefore
leads to a marked damping of the wave. This is evident from Fig. 64. When
the discontinuity is formed, the highest part of the wave profile is cut off.
In the course of time, as the profile is bent over, its height becomes less, and
the profile is "smoothed" to one of smaller amplitude, i.e. the wave is damped.
Fig. 64
It is clear from the above that discontinuities must ultimately be formed
in every simple wave which contains regions where the density decreases in the
direction of propagation. The only case where discontinuities do not occur
is a wave in which the density increases monotonically in the direction of
propagation (such, for example, is the wave formed when a piston moves
out of an infinite pipe filled with gas; see the Problems at the end of this
section).
Although the wave is no longer a simple one when a discontinuity has been
formed, the time and place of formation of the discontinuity can be deter
mined analytically. We have seen that the occurrence of discontinuities is
mathematically due to the fact that, in a simple wave, the quantities/), p and v
become manyvalued functions of x (for given t) at times greater than a
370 Onedimensional Gas Flow §94
certain definite value t , whereas for t < to they are onevalued functions.
The time *o is the time of formation of the discontinuity. It is evident from
geometrical considerations that, at the instant t , the curve giving, say, v
as a function of * becomes vertical at some point x = xo, which is the point
where the function is subsequently manyvalued. Analytically, this means
that the derivative (8vjdx) t becomes infinite, and (Bxjdv) t becomes zero. It
is also clear that, at the instant to, the curve v = v(x) must lie on both
sides of the vertical tangent, since otherwise v(x) would already be many
valued. In other words, the point x = x must be, not an extremum of the
function x(v), but a point of inflexion, and therefore the second derivative
(d 2 xldv% must also vanish. Thus the place and time of formation of the
shock wave are determined by the simultaneous equations
(dx/dv) = 0, (82x[dv2) t = 0. (94.12)
For a perfect gas these equations are
* = 2/»/(y+l), /» = 0, (94.13)
where f(v) is the function appearing in the general solution (94.8).
These conditions require modification if the simple wave adjoins a gas at
rest and the shock wave is formed at the boundary. Here also the curve
v = v{x) must become vertical, i.e. the derivative (dx[8v) t must vanish, at
the time when the discontinuity occurs. The second derivative, however,
need not vanish; the second condition here is simply that the velocity is
zero at the boundary of the gas at rest, so that (dxjdv) t = for v = 0. From
this condition we can obtain explicit expressions for the time and place of
formation of the discontinuity. Differentiating (94.5), we obtain
*= /'(0)K *= ±c *+/(0), (94.14)
where ao is the value, for v = 0, of the quantity a defined by formula (95.2).
For a perfect gas
*= 2/'(0)/(y+l). (94.15)
PROBLEMS
Problem 1 . A perfect gas is in a semiinfinite cylindrical pipe (x > 0) terminated by a piston.
At time t — the piston begins to move with a uniformly accelerated velocity U = ±at.
Determine the resulting flow.
Solution. If the piston moves out of the pipe (U = — at), the result is a simple rare
faction wave, whose forward front is propagated to the right, through gas at rest, with
velocity c ; in the region x > c t the gas is at rest. At the surface of the piston, the gas and
the piston must have the same velocity, i.e. we must have v = —at for x = —%at 2 (t > 0).
This condition gives for the function /(*>) in (94.8)
f(at) = cot + %yat*.
Hence we have
*>o+Ky +1X1* =/(*>)
= covJa+\yv 2 la t
§94
Onedimensional travelling waves
371
whence
v = [co+Kr + iKI/y V(h+Ky+ iKI 2 2«y(<;o**)}/y. (l)
This formula gives the change in velocity over the region between the piston and the forward
front x = c t of the wave (Fig. 65a) during the time interval t = to t = 2c l(y—l)a.
The gas velocity is everywhere to the left, like that of the piston, and decreases monotonically
in magnitude in the positive ^direction; the density and pressure increase monotonically in
that direction. For t > 2cJ(y—l)a, the inequality (94.10) does not hold for the piston
velocity, and so the gas can no longer follow the piston. A vacuum is then formed in a region
adjoining the piston, beyond which the gas velocity decreases from — 2c /(y— 1) to zero
according to formula (1).
If the piston moves into the pipe (U = at), a simple compression wave is formed; the
corresponding solution is obtained by merely changing the sign of a in (1) (Fig. 65b). It is
valid, however, only until a shock wave is formed ; the time when this happens is determined
from formula (94.15), and is
t = 2c Q ja{y+\).
Problem 2. The same as Problem 1, but for the case where the piston moves in any
manner.
Solution. Let the piston begin to move at time t = according to the law x — X(t)
(with X(0) = 0); its velocity is U = X'{t). The boundary condition on the piston (v = U
for x = X) gives v = X'(t),f(v) = X(t)t[c +Uy + l)X'it)]. If we now regard t as a para
meter, these two equations determine the function f(v) in parametric form. Denoting the
parameter by t, we can write the solution as
v = X\r), x = X{T) + {tT)[c +\{y+\)X\T)l
(1)
which determines, in parametric form, the required function v(t, x) in the simple wave which
is caused by the motion of the piston.
Problem 3. Determine the time and place of formation of the shock wave when the piston
(Problem 1) moves according to the law U = at n (m > 0).
Solution. If a < 0, i.e. the piston moves out of the pipe, a simple rarefaction wave
results, in which no shock wave is formed. We therefore assume that a > 0, i.e. the piston
moves into the pipe, causing a simple compression wave.
372 Onedimensional Gas Flow §95
When the function v(x, t) is given by the parametric formulae (1) (Problem 2), and
X = aT n+1 jn + l, the time and place of formation of the shock wave are given by the
equations
(8x \
— I = CO + ^lfl W (y+l)lfl T «[ y l +W ( y+ l)] = 0,
/ 8 2 x \ ' '
i^J = itTn2an(nl)(y+l)^anrni[yl+n(y+l)] = 0,
where the second equation must be replaced by t = if we are concerned with the formation
of a shock wave at the forward front of the simple wave.
For n = 1 we find t = 0, t = 2c /a(y + l), i.e. the shock wave is formed at the forward
front at a finite time after the motion begins, in accordance with the results of Problem 1.
For n < 1, the derivative 8x/8t is of varying sign (and therefore the function v(x) for given
t is manyvalued) for any t > 0. This means that a shock wave is formed at the piston as
soon as it begins to move.
For n > 1 the shock wave is formed, not at the forward front of the simple wave, but at
some intermediate point given by (1). Having determined r and t from (1), we can then
find the place of formation of the discontinuity from (1) of Problem 2. The result is
\ a I y+llnr J
„ /2co\ 1/re r y nll 1
* = 2cq\ 1 — .
\ a J Ly+1 »+ 1 J (nl)<»i>/»[yl + n(y +l)]i/»
§95. Formation of discontinuities in a sound wave
A travelling plane sound wave, being an exact solution of the equations of
motion, is also a simple wave. We can use the general results obtained in §94
to derive some properties of sound waves of small amplitude in the second
approximation (the first approximation being that which gives the ordinary
linear wave equation).
We must notice first of all that a discontinuity must ultimately appear in
each wavelength of a sound wave. This leads to a very marked damping of
the wave, as shown in §94. It must be remarked, however, that this happens
only for a sufficiently strong sound wave ; a weak sound wave is damped by
the usual effects of viscosity and thermal conduction before the effects of
higher order in the amplitude can develop.
The distortion of the wave profile has another effect also. If the wave
is purely harmonic at some instant, it ceases to be so at later instants, on
account of the change in shape of the profile. The motion, however, remains
periodic, with the same period as before. When the wave is expanded in a
Fourier series, terms with frequencies nco (n being integral and co being the
fundamental frequency) appear, as well as that with frequency co. Thus the
distortion of the profile as the sound wave is propagated may be regarded as
the appearance in it of higher harmonics in addition to the fundamental
frequency.
§95 Formation of discontinuities in a sound wave 373
The velocity u of points in the wave profile (the wave being propagated in
the positive ^direction) is obtained, in the first approximation, by putting in
(94.11) v = 0, i.e. u = cq, corresponding to the propagation of the wave
with no change in its profile. In the next approximation we have
u = co + p' du/dpo = co+(du/dp )povjco,
or, using the expression (92.10) for the derivative dujdp,
u = co + ao^, (95.1)
where we have put for brevity
a = (c*l2V*)(&Vldp)s. (95.2)
For a perfect gas, a = ^(y+l), and formula (95.1) agrees with the exact
formula (see (94.8)) for the velocity u.
In the general case of arbitrary amplitude, the wave is no longer simple
after the discontinuities have appeared. A wave of small amplitude, however,
is still simple in the second approximation even when discontinuities are
present. This can be seen as follows. The changes in velocity, pressure and
specific volume in a shock wave are related by #2—^1 = VKP2— Pi)(Vi— V2)].
The change in the velocity v over a segment of the #axis in a simple wave is
v 2 v! = J ^(8V/dp)dp.
Pi
A simple calculation, using an expansion in series, shows that these two
expressions differ only by terms of the third order (it must be borne in mind
that the change in entropy at a discontinuity is of the third order of smallness,
while in a simple wave the entropy is constant). Hence it follows that, as far
as terms of the second order, a sound wave on either side of a discontinuity in
it remains simple, and the appropriate boundary condition is satisfied at the
discontinuity itself. In higher approximations this is no longer true, on
account of the appearance of waves reflected from the surface of discontinuity.
Let us now derive the condition which determines the location of the dis
continuities in a travelling sound wave (again in the second approximation).
Let u be the velocity of the discontinuity relative to a fixed coordinate
system, and v\, v% the velocities of the gases on each side of it. Then the
condition that the mass flux is continuous is pi(^i — u) = ^2(^2 — u), whence
u — {p\v\ — p2P2)\(pi — p%). As far as the secondorder terms, this is equal to
the derivative d(pv)ldp at the point where v is equal to \{vi + V2) :
u = [d(pv)/dp] v = i(Vl+Vt) .
Since, in a simple wave, d{pv)jdp = v + c, we have, by (95.1),
u = co + i<x(^i+^2) (95.3)
From this we can obtain the following simple geometrical condition which
determines the position of the shock wave. In Fig. 66 the curve shows the
velocity profile corresponding to the simple wave ; let ae be the discontinuity.
13
374 Onedimensional Gas Flow §95
The difference of the shaded areas abc and cde is the integral
{x~xq)6v
taken along the curve abcde. In the course of time, the wave profile moves ;
v 2
*0
Fig. 66
let us calculate the time derivative of the above integral. Since the velocity
dx/dt of points in the wave profile is given by formula (95.1), and the velocity
dxojdt of the discontinuity by (95.3), we have
dt
V* V% Vt
J (x— xo)d?; = ol{ J vdv—%(vi+V2) J dv} = 0;
in differentiating the integral, we must notice that, although the limits of
integration v\ and V2 also vary with time, x—xo always vanishes at the limits,
and so we need only differentiate the integrand.
Thus the integral J* (x— #o)dz> remains constant in time. Since it is zero at
the instant when the shock wave is formed (the points a and e then coin
ciding), it follows that we always have
I (x— #o)dz> = 0.
(95.4)
abcde
Geometrically this means that the areas abc and cde are equal, a condition
which determines the position of the discontinuity.
Let us consider a single one dimensional compression pulse, in which a
shock wave has already been formed, and ascertain how this shock will finally
be damped. By so doing, we also find the law of damping of any plane shock
wave after it has been propagated for a sufficiently long time.
In the later stages of its propagation, a sound pulse containing a shock
wave will have a triangular velocity profile. Let the profile be given at some
instant (which we take as t = 0) by the triangle ABC (Fig. 67a). If the
points in this profile move with the velocities (95.1), we obtain after time /
§95
Formation of discontinuities in a sound wave
375
a profile A'B'C (Fig. 67b). In reality, the discontinuity moves to E, and
the actual profile will be A'DE. The areas DB'F and C'FE are equal, by
(95.4), and therefore the area A'DE of the new profile is equal to the area
ABC of the original profile. Let / be the length of the sound pulse at time t,
and Av the velocity discontinuity in the shock wave. During time t> the point
B moves a distance cutAvo relative to C; the tangent of the angle B'A'C is
therefore A^ /(/o + cutAvo), and we obtain the condition of equal areas ABC
and A'DE in the form
whence
IqAvq = Z 2 A^ /(/o + a/A^o),
/ = /oVU + aA^/A)],
Av = Avol^[l + oLAvotjlo].
(95.5)
(a)
Av n
Fig. 67
For t > oo the intensity of the shock wave diminishes asymptotically with
time as \\*Jt (or, what is the same thing, with distance as Ify/x). The
total energy of a travelling sound pulse (per unit area of its front) is
E = po j v 2 dx = Eol V[l + aA*; tj%
(95.6)
where Eo is the energy at time t = 0. For t > oo the energy also tends to
zero as Ijy/t.
If we have a spherical outgoing sound wave, any small section of it can be
regarded as plane at sufficiently large distances r from the origin. The
velocity of any point in the wave profile is then given by formula (95.1). If,
however, we wish to use this formula to follow the motion of any point in the
wave profile over long intervals of time, we must take into account the fact
that the amplitude of a spherical wave falls off inversely as the distance r,
even in the first approximation. This means that, at any given point in the
profile, v is not constant, as it is for a plane wave, but decreases as 1/r. If
376 Onedimensional Gas Flow §95
vq is the value of v (for a given point in the profile) at a (large) distance yq,
we can put v = vorofr. Thus the velocity u of points in the wave profile is
u — co+(x.voro[r. The first term is the ordinary velocity of sound, and cor
responds to movement of the wave without change in the shape of the profile
(apart from the general decrease of the amplitude as \fr). The second term
results in a distortion of the profile. The amount 8r of this additional move
ment of points in the profile during a time t = (r — yq)Jc is obtained by multi
plying by drjCQ and integrating from yq to r ; this gives
8y = (ccvoYolco)log(YJYo). (95.7)
Thus the distortion of the profile of a spherical wave increases as the logarithm
of the distance, i.e. much more slowly than for a plane wave, where the dis
tortion Sx increases as the distance x traversed by the wave.
Fig. 68
The distortion of the profile ultimately leads to the formation of dis
continuities in it. Let us consider shock waves formed in a single spherical
sound pulse which has reached a large distance from the source (the origin).
The spherical case is distinguished from the plane case primarily by the fact
that the region of compression must be followed by a region of rarefaction ;
the excess pressure and the velocity of the gas particles in the wave must both
change sign (see §69). The distortion of the profile results ultimately in the
formation of two shock waves: one in the region of compression, and the
other in the region of rarefaction (Fig. 68). f In the leading shock wave, the
pressure increases discontinuously, then gradually decreases into a rarefac
tion, then again increases discontinuously in the second shock (but not to
its unperturbed value, which is reached only asymptotically behind this
shock).
The manner of the final damping of the shock waves with time (or, what
is the same thing, with the distance r from the source) is easily found in
exactly the same way as for the plane case discussed above. Using the result
(95.7), we find that, at sufficiently large distances, the thickness / of the sound
f It should be mentioned that, since there is always ordinary damping (due to viscosity and
thermal conduction) when sound is propagated in the gas, the slowness of the distortion in a spherical
wave may have the result that it is damped before discontinuities can be formed.
§95
Formation of discontinuities in a sound wave
377
pulse (the distance between the two discontinuities) increases as log*(r/a),
instead of as \/x for the plane case ; a is some constant length. The intensity
of the leading shock wave is damped according to rAv ~ log~ ¥ (r/a), or
A© ~ lfr \ogi(r/a). (95.8)
Finally, let us consider the cylindrical case. The general decrease in the
amplitude of an outgoing sound wave occurs in inverse proportion to yV,
where r is the distance from the axis. Repeating the arguments given for the
spherical case, we now find the velocity u of points in the wave profile to be
u = co + a.vo\/(rolr), and so the displacement Sr of points in the profile,
between ro and r is
Sr = 2a(volco)\/ro(<\/r ^/r ).
(95.9)
Fig. 69
The cylindrical propagation of a compression pulse must be accompanied,
as in the spherical case, by a rarefaction of the gas behind the compression.
Two shock waves must therefore be formed in this case also. By the same
method, we find the ultimate law of increase of the thickness of the pulse :
/ ~ r% and the ultimate law of damping of the intensity of the shock wave :
y/rAv ~ r _i , or
Av ~ rS. (95.10)
The formation of discontinuities in a sound wave is an example of the
spontaneous occurrence of shock waves in the absence of any singularity in
the external conditions of the flow. It must be emphasised that, although a
shock wave can appear spontaneously at a particular instant, it cannot dis
appear in the same manner. Once formed, a shock wave decays only asymp
totically as the time becomes infinite.
PROBLEMS
Problem 1. At the initial instant, the wave profile consists of an infinite series of "teeth",
as shown in Fig. 69. Determine how the profile and energy of the wave change with time.
Solution. It is evident that, at subsequent instants, the wave profile will be of the same
form, with / unchanged but the height vt less than v . Let us consider one "tooth" : at time
t — 0, the ordinate through the point where v — vt cuts off a part vtljv of the base of the
triangle. During a time t, this point moves forward a distance ocvtf. The condition that the
base of the triangle is unchanged in length is vtl lv +a.tvt = l , whence vt — *> /(l +ctv t/l ).
As t *■ oo, the wave amplitude diminishes as \jt. The energy is E = E /(l +<zv t]l o y, i.e. it
diminishes as ljt 2 for t > co.
378 Onedimensional Gas Flow §96
Problem 2. Determine the intensity of the second harmonic formed by the distortion of
the profile of a monochromatic spherical wave.
Solution. Writing the wave in the form rv — A cos(kr—tot), we can allow for the distor
tion, in the first approximation, by adding Sr to r on the righthand side of this equation, and
expanding in powers of Sr. This gives, by (95.7),
rv — A cos(kr—cQt) — (oLk/2co)A 2 log(r/ro) sin 2(kr— cot);
here r must be taken as a distance at which the wave can still be regarded, with sufficient
accuracy, as strictly monochromatic. The second term in this formula is the second harmonic
in the spectral resolution of the wave. Its total (time average) intensity I % is
h = (aP&l&rcoSpo) lo g 2(r/r )/i 2 ,
where 1^ = IttCqPqA 2 is the intensity of the first harmonic.
§96. Characteristics
The definition of characteristics, given in §79, as lines along which small
disturbances are propagated (in the approximation of geometrical acoustics) is
of general validity, and is not restricted to the plane steady supersonic flow
discussed in §79.
For onedimensional nonsteady flow, we can introduce the characteristics
as lines in the artplane whose slope dxjdt is equal to the velocity of propaga
tion of small disturbances relative to a fixed coordinate system. Disturbances
propagated relative to the gas with the velocity of sound, in the positive or
negative ^direction, move relative to the fixed coordinate system with
velocity v±c. The differential equations of the two families of characteristics,
which we shall call C + and C_, are accordingly
(dx/dt)+ = v + c, (dx/dt) = vc. (96.1)
Disturbances transmitted with the gas are propagated in the atfplane along
characteristics belonging to a third family Co, for which
(d*/d*)o = v. (96.2)
These are just the "streamlines" in the atfplane; cf. the end of §79.f It
should be emphasised that, for characteristics to exist, it is no longer necessary
for the gas flow to be supersonic. The "directional" propagation of distur
bances, as evidenced by the characteristics, is here simply due to the causal
relation between the motions at successive instants.
As an example, let us consider the characteristics of a simple wave. For a
wave propagated in the positive ^direction we have, by (94.5), x = t(v + c) +
+f(v). Differentiating this relation, we have
dx = (v + c)dt + [t + tc'(v)+f\v)]dv.
Along a characteristic C + , we have dx = {v + c)dt ; comparing the two equa
tions, we find that along such a characteristic [t + tc'(v)+f'(v)]dv = 0. The
t The same equations (96.1) and (96.2) determine the characteristics for nonsteady spherically
symmetrical flow, if x is replaced by the radial coordinate r (the characteristics now being lines in
the riplane).
§96
Characteristics
379
expression in brackets cannot vanish identically, and therefore dv = 0, i.e.
v = constant. Thus we conclude that, along any characteristic C + , the velo
city is constant, and therefore so are all other quantities. The same property
holds for the characteristics C in a wave propagated to the left. We shall see
in §97 that this is no accident, but is a mathematical consequence of the nature
of simple waves.
From this property of the characteristics C + for a simple wave, we can in
turn conclude that they are a family of straight lines in the atfplane; the
velocity is constant along the lines x = t[v + c(v)] +f(v) (94.5). In particular,
for a similarity rarefaction wave (a simple wave with f(v) = 0), these lines
form a pencil through the origin in the artplane. For this reason, a similarity
simple wave is sometimes said to be centred.
Fig. 70
Fig. 70 shows the family of characteristics C + for the simple rarefaction
wave formed when a piston moves out of a pipe with acceleration. It is a family
of diverging straight lines, which begin from the curve x = X(t) giving the
motion of the piston. To the right of the characteristic x = c$t lies a region
of gas at rest, where the characteristics become parallel.
Fig. 71 is a similar diagram for the simple compression wave formed when a
piston moves into a pipe with acceleration. In this case the characteristics are
converging straight lines, which eventually intersect. Since every charac
teristic has a constant value of v, their intersection shows that the function
v(x, t) is manyvalued, which is physically meaningless. This is the geo
metrical interpretation of the result obtained in §94: a simple compression
wave cannot exist indefinitely, and a shock wave must be formed in it. The
geometrical interpretation of the conditions (94.12), which determine the
time and place of formation of the shock wave, is as follows. The intersecting
family of rectilinear characteristics has an envelope, which, for a certain
least value of t, has a cusp ; this gives the instant at which manyvaluedness
first occurs. Every point in the region between the two branches of the en
velope is on three characteristics C + . If the equations of the characteristics
380
Onedimensional Gas Flow
§96
are given in the parametric form x = x(v), t = t(v), the position of the cusp
is given by equations (94.12).f
We shall now indicate briefly how the physical definition, given above, of
the characteristics as lines along which disturbances are propagated corre
sponds to the mathematical sense of the word in the theory of partial diff
erential equations. Let us consider a partial differential equation of the form
dU 8U d 2 d>
A— + IB —  + C—  + D = 0,
dx 2 dx dt 8t 2
(96.3)
Envelope
Fig. 71
which is linear in the second derivatives; the coefficients A, B, C, D can be
any functions, both of the independent variables x, t and of the unknown
function ^ and its first derivatives. $ Equation (96.3) is of the elliptic type if
B 2 — AC < everywhere, and of the hyperbolic type if B 2 — AC > 0. In
the latter case, the equation
Adt 2 2Bdxdt+Cdx 2 = 0,
(96.4)
or
dx/dt =[B± V(B 2 AC)]/C, (96.5)
determines two families of curves in the xtplane, the characteristics (for a
given solution <f>(x, t) of equation (96.3)). We may point out that, if the co
efficients A, B, C are functions only of x and t, then the characteristics are
independent of the particular solution </>.
Let a given flow correspond to some solution <f> — <f>o(x, t) of equation
(96.3), and let a small perturbation <£i be applied to it. We assume that this
perturbation satisfies the conditions for geometrical acoustics to be valid:
it does not greatly affect the flow (<£i and its first derivatives are small), but
f The particular case where the shock wave occurs at the boundary of the gas at rest corresponds
to that where one branch of the envelope is part of the characteristic x = c t.
J The velocity potential satisfies an equation of this form in onedimensional nonsteady flow.
§97 Riemann invariants 381
varies considerably over small distances (the second derivatives of <f>i are
relatively large). Putting in equation (96.3) <f> = fo + fa, we then obtain
for <f>\ the equation
d^d>! dUi dtfa
dx 2 dx dt dt 2
with <j> = <f> in the coefficients A, B, C. Following the method used in
changing from wave optics to geometrical optics, we write <f>i = ae^ t where
the function iff (the eikonal) is large, and obtain
^)Wi^ + C (^) 2 = 0. (96.6)
\ 8x / dx dt \dtl
The equation of ray propagation in geometrical acoustics is obtained by
equating dxjdt to the group velocity: dx[dt = dcojdk, where k = tyjdx,
co^difjjdt. Differentiating the relation Ak?2Bkco + Cco 2 = 0, we
obtain dxjdt = (Ba)Ak)fcCa>Bk), and, eliminating k\oi by the same
relation, we again arrive at equation (96.5).
PROBLEM
Find the equation of the second family of characteristics in a centred simple wave.
Solution. In a centred simple wave propagated into gas at rest to the right of it, we have
xjt = v+c = c +Ur+l)v. The characteristics C + form the pencil * = constant Xt. The
characteristics C_, on the other hand, are determined by the equation
dx 3 — y x 4
— = V — C = 7 7^0
dt y+1 t y+l
Integrating, we find
2 y+l ft \<37>/<r+D
x = cot H com — 1 ,
y— 1 y—\ \ to /
where the constant of integration has been chosen so that the characteristic C_ passes through
the point * = c t , t = 1 on the characteristic C+ (x = c t) which is the boundary between
the simple wave and the region at rest.
The "streamlines" in the xfplane are given by the equation
dx 2
whence
dx 2 / X \
dt y+l\t J
2 y+l It \ 2/( t +1)
x — cot \ co^ol — 1
1 / * '
coto\ —
1 \ to,
y—\ y—\ \ to I
§97. Riemann invariants
An arbitrary small disturbance is in general propagated along all three
characteristics (C + , C, Co) leaving a given point in the atfplane. However,
an arbitrary disturbance can be separated into parts each of which is pro
pagated along only one characteristic.
382 Onedimensional Gas Flow §97
Let us first consider isentropic gas flow. We write the equation of con
tinuity and Euler's equation in the form
dp dp dv
~ + v— + pc*~ = 0,
dt dx dx
dv dv 1 dp
— + v— + — = 0;
dt dx p dx
in the equation of continuity we have replaced the derivatives of the density
by those of the pressure, using the formulae
dp
dt \dp/ s
Dividing the first equation by ± pc and adding it to the second, we obtain
dv 1 dp (dv 1 dp
— ± + — ±
dt pc dt \ dx pc dx
We now introduce as new unknown functions
I dp \ dp 1 dp dp 1 dp
~ \dp) s dt~ c 2 dt' dx ~ c 2 dx
)(v±c) = 0. (97.1)
J+ = v+jdplpc, J = vjdpl P c, (97.2)
which are called Riemann invariants. It should be remembered that, in
isentropic flow, p and c are definite functions of p, and the integrals on the
righthand sides are therefore definite functions. For a perfect gas
J + = v + 2cl(yl), J = v2cl(yl). (97.3)
In terms of these quantities, the equations of motion take the simple form
Yd d 1 X d d 1
fc+^d^ ' [» + <"w° < 97  4)
The differential operators acting on J + and /_ are just the operators of
differentiation along the characteristics C + and C in the atfplane. Thus we
see that / + and /_ remain constant along each characteristic C + or C re
spectively. We can also say that small perturbations of J + are propagated
only along the characteristics C + , and those of/ only along C—
In the general case of anisentropic flow, the equations (97.1) cannot be
written in the form (97.4), since dpfpc is not a perfect differential. These
equations, however, still permit the separation of perturbations propagated
along characteristics of only one family. For such perturbations are those
of the form Sv ± Spjpc, where 8v and Sp are arbitrary small perturbations
of the velocity and pressure. In order to obtain a complete system of equa
tions of motion, the equations (97.1) must be supplemented by the adiabatic
equation
Yd dl
[« + "d' ' (97  5)
which shows that perturbations 8s are propagated along the characteristics Co
§97
Riemann invariants
383
An arbitrary small perturbation can always be separated into independent
parts of the three kinds mentioned.
A comparison with formula (94.4) shows that the Riemann invariants (97.2)
are the quantities which, in simple waves, are constant throughout the region
of the flow at all times : / is constant in a simple wave propagated to the right,
and J + in one travelling to the left. Mathematically, this is the fundamental
property of simple waves, from which follows, in particular, the property
mentioned in §96: one family of characteristics consists of straight lines. For
example,l et the wave be propagated to the right. Each characteristic C + has
a constant value of / + and, furthermore, a constant value of/, which value is
the same everywhere. Since both / + and /_ are constant, it follows that y
and/» are constant (and therefore so are all the other quantities), and we obtain
the property of the characteristics C + deduced in §96, which in turn shows
that they are straight lines.
1 Simple wave
2 Constant flow
Fig. 72
If the flow in two adjoining regions of the artplane is described by two
analytically different solutions of the equations of motion, then the boundary
between the regions is a characteristic. For this boundary is a discontinuity
in the derivatives of some quantity, i.e. it is a weak discontinuity, and there
fore must necessarily coincide with some characteristic.
The following property of simple waves is of great importance in the theory
of onedimensional isentropic flow. The flow in a region adjoining a region of
constant flow (in which v = constant, p = constant) must be a simple wave.
This statement is very easily proved. Let the region 1 in the atfplane be
bounded on the right by a region (2) of constant flow (Fig. 72). Both in
variants J + and/ are evidently constant in the latter region, and both families
of characteristics are straight lines. The boundary between the two regions
is a characteristic C + , and the lines C + in one region do not enter the other
region. The characteristics C pass continuously from one region to the other,
and carry the constant value of /_ into region 1 from region 2. Thus J is
constant throughout region 1 also, so that the flow in the latter is a simple
wave.
384 Onedimensional Gas Flow §97
The ability of characteristics to "transmit" constant values of certain
quantities throws some light on the general problem of initial and boundary
conditions for the equations of fluid dynamics. In particular cases of
physical interest, there is usually no doubt about the choice of these condi
tions, which is dictated by physical considerations. In more complex cases,
however, mathematical considerations based on the general properties of
characteristics may be useful.
1 quantity 2 quantities Iquantity 1 quantity 2 quantities 2 quantities
Ia
Fig. 73
For definiteness, we shall discuss a onedimensional isentropic gas flow.
Mathematically, a problem of gas dynamics usually amounts to the deter
mination of two unknown functions (for instance, v and p) in a region of the
atfplane lying between two given curves {OA and OB in Fig. 73a), on which
the boundary conditions are known. The problem is to find how many
quantities can take given values on these curves. In this respect it is very
important to know how each curve is situated relative to the directions (shown
by arrows in Fig. 73) of the two characteristics C + and C_ leavingf each
point of it. Two cases can occur: either both characteristics lie on the same
side of the curve, or they do not. In Fig. 73 a, the curve OA belongs to the
first case and the curve OB to the second. It is clear that, for a complete
determination of the unknown functions in the region AOB, the values of
two quantities must be given on the curve OA (e.g. the two invariants / +
and /_), and those of only one quantity on OB. For the values of the second
quantity are "transmitted" to the curve OB from the curve OA by the charac
teristics of the corresponding family, and therefore cannot be given arbi
trarily.! Similarly, Figs. 73b and c show cases where one and two quantities
respectively are given on each bounding curve.
It should also be mentioned that, if the bounding curve coincides with a
characteristic, two independent quantities cannot be specified on it, since
t In the artplane, the characteristics leaving a given point are those which go in the direction of
t increasing.
J An example of this case may be given as an illustration : the gas flow when a piston moves into or
out of an infinite pipe. Here we are concerned with finding a solution of the equations of gas dynamics
in the region of the artplane lying between two lines, the positive a>axis and the line x = X(t) which
gives the movement of the piston (Figs. 70, 71). On the first line the values of two quantities are
given (the initial conditions v = 0, p = p for t = 0), and on the second line those of one quantity
(v = u, where u(t) is the velocity of the piston).
§97 Riemann invariants 385
their values are related by the condition that the corresponding Riemann
invariant is constant.
The problem of specifying boundary conditions for the general case of
anisentropic flow can be discussed in an entirely similar manner.
Finally, we may make the following remark. We have everywhere above
spoken of the characteristics of onedimensional flow as lines in the xt plane.
The characteristics can, however, also be defined in the plane of any two
variables describing the flow. For example, we can consider the characteris
tics in the wplane. For isentropic flow, the equations of these characteristics
are given simply by / + = constant, J = constant, with various constants
on the right; we call these characteristics V + and T_. For a perfect gas these
are, by (97.3), two families of parallel lines (Fig. 74).
Fig. 74
It should be noted that these characteristics are entirely determined by
the properties of the gas, and do not depend on any particular solution of the
equations of motion. This is because the equation of isentropic flow in the
variables v, c is (as we shall see in §98) a linear secondorder partial differen
tial equation with coefficients which depend only on the independent vari
ables.
The characteristics in the xt and vc planes are transformations of one
another involving the particular solution of the equations of motion. The
transformation need not be onetoone, however. In particular, only one
characteristic in the wplane corresponds to a given simple wave, and all the
characteristics in the atfplane are transformed into it. For a wave travelling
to the right (e.g.), it is one of the characteristics T_; the characteristics C_
are transformed into the line T, and the characteristics C + into its various
points.
PROBLEM
Find the general solution of the equations of onedimensional isentropic flow of a perfect
gas with y = 3.
Solution. For y = 3 we have J±= v±c, and equations (97.4) have the general integral
x = (v + c)t+fi(v + c),
x = (v — c)t+fz(v — c),
where / x and/ 2 are arbitrary functions. These two equations implicitly determine the required
3 86 Onedimensional Gas Flow
§98
functions v(x, t) and c(x, t), and therefore all other quantities. We may say that, in this case,
the two quantities v±c are propagated independently as two simple waves which do not
interact.
§98. Arbitrary onedimensional gas flow
Let us now consider the general problem of arbitrary onedimensional
isentropic gas flow (without shock waves). We shall first show that this
problem can be reduced to the solution of a linear differential equation.
Any onedimensional flow (i.e. a flow depending on only one spatial coor
dinate) must be a potential flow, since any function v(x, t) can be written
as a derivative: v(x, t) = dcf>{x, t)/dx. We can therefore use, as a first integral
of Euler's equation, Bernoulli's equation (9.3): d<j>jdt+\v 2 + w = 0. From
this, we find the differential
8<f> dd>
d<f> = —dx + — dt
8x dt
= vdx(\v 2 + w)dt.
Here the independent variables are x and t; we now change to the independent
variables v and w. To do so, we use Legendre's transformation; putting
d<j> = d(xv)xdvd[t(w + ±v2)] + td(w+lv 2 )
and replacing <f> by a new auxiliary function
X = <l> — xv + t(w+%v 2 ),
we obtain
dx = xdv + td(zo+%v 2 ) = tdw + (vtx)dv>
where x is regarded as a function of v and w. Comparing this relation with
the equation d* = (d x ldw)dw + (d x ldv)dv, we have t = d x [dw, vtx
= dx/dv, or
t = dx/dw, x = vdx/dwdxldv. (98.1)
If the function xfa, «>) is known, these formulae determine v and to as
functions of the coordinate x and the time t.
We now derive an equation for X  To do so, we start from the equation
of continuity, which has not yet been used:
dp 8 dp dp dv
— + —{pv) = — + v— + p— = 0.
dt dx K dt dx H dx
We transform this equation to one in terms of the variables v, to. Writing
the partial derivatives as Jacobians, we have
d(p,x) d(t,p) d(t,v)
h v f p = 0.
d(t,x) d(t,x) H d(t,x)
§98 Arbitrary onedimensional gas flow 387
or, multiplying by d(t, x)jd(w, v),
d(p,x) 8(t,p) d(t,v)
1 V h p = U.
d(w,v) d(w,v) d(zv,v)
To expand these Jacobians we must use the following result. According to
the equation of state of the gas, p is a function of any two other independent
thermodynamic quantities; for example, we may regard p as a function of w
and s. If s = constant, we have simply p = p(w), and the density is inde
pendent of v. Expanding the Jacobians, we therefore have
dp 8x dp dt 8t
dw dv dw dv dw
Substituting here the expressions (98.1) for t and x, we obtain
p dw \ dw dv 2 ) dw 2
Ifs = constant, we have dw = dp/p, whence dwjdp = 1/p. We can therefore
write dp/dro = (dp/d»(dp/dw) = p\c 2 . We finally have for x the equation
c *x_*x + e L= (98 . 2)
8w 2 dv 2 dw
here the velocity of sound c is to be regarded as a function of w. The problem
of integrating the nonlinear equations of motion has thus been reduced
to that of solving a linear equation.
Let us apply this result to the case of a perfect gas. We have c 2 = (y \)w,
and the fundamental equation (98.2) becomes
(y _ 1)K ^_£^ + ^ = 0. (98.3)
v/ ; dw 2 dv 2 dw
This equation has an elementary general integral if (3y)/(y 1) is an even
integer :
(3_ y )/( y _l) = 2«, or y = (3 + 2«)/(2«+l), w = 0,1,2,.... (98.4)
This condition is satisfied by monatomic (y = I, n = 1) and diatomic
(y = 7 1 n = 2) gases. Expressing y in terms of n, we can rewrite (98.3) as
2 ^_f* + ^ = . (98.5)
2n+l dw 2 dv 2 dw
We denote by xn a function which satisfies this equation for a given n
For the function xo we have
d 2 X o d 2 X o d X o
2w — H = U.
dw 2 dv 2 dw
388 Onedimensional Gas Flow §98
Introducing in place of to the variable u = \/(2w), we obtain
This is just the ordinary wave equation, whose general solution is
Xo = fi(u+v)+f 2 (uv),
/i and/ 2 being arbitrary functions. Thus
Xo = /i[V(2«>) +v] +f 2 [V(2to)v]. (98.6)
We shall now show that, if the function X n is known, the function X n+i
can be obtained by differentiation. For, differentiating equation (98.5) with
respect to w, we easily find
2 & 2 /8 Xn \ 2n + 3 d / d Xn \ d 2 / d x
^f+^T7f TT^ =0.
2n+l dw 2 \ dw J 2n+l dw \ dw
I dv 2 \dw)
Putting v = v'V[(2n+ 1)/(2« + 3)], we have for d Xn \dw the equation
2n + 3 dw 2 \ dw J dw\dw) 8v' 2 \ dw / '
which is equation (98.5) for the function X n+\ (w, v'). Thus we conclude that
( >\ d t ^ d I /2«+l\
Xn+1 («,« ) = — *,(»,,,) = — Xn (^y— )• (98.7)
Using this formula n times and taking Xo from (98.6), we find that the
general solution of equation (98.5) is
Qn
X = ^^^ 2 ( 2n+1 ^ +v 'i + MV[2(2n+l)w]v]},
or
3" 1 f F 1 [V[2{2n+l)w]+v] + F 2 [V[2(2n+l)w]v])
x = ^( ^ ~ J — i }' ( 98  8 )
where Fi and F 2 are again two arbitrary functions.
If we express w in terms of the velocity of sound by w = c 2 l(yl)
= %{2n + l)c 2 , the solution (98.8) becomes
* = ay ]H c + ^ti) + h2^t)) < 98  9 )
The expressions c± v((2n+ 1) = c± £(y l)v which are the arguments of the
arbitrary functions are just the Riemann invariants (97.3), which are constant
along the characteristics.
§98 Arbitrary onedimensional gas flow 389
In applications it is often necessary to calculate the values of the function
x{v, c) on a characteristic. The following formulaf is useful for this purpose :
with ±vj(2n+l) = c + a (a being an arbitrary constant).
Let us now ascertain the relation between the general solution just found
and the solution of the equations of gas dynamics which describes a simple
wave. The latter is distinguished by the property that in it v is a definite
function of to: v = v(w), and therefore the Jacobian A = d(v, w)jd(x, t)
vanishes identically. In transforming to the variables v and zo, however, we
divided the equation of motion by this Jacobian, and the solution for which
A = is therefore "lost". Thus a simple wave cannot be directly obtained
from the general integral of the equations of motion, but is a special integral
of these equations.
To understand the nature of this special integral, we must observe that it
can be obtained from the general integral by a certain passage to a limit,
which is closely related to the physical significance of the characteristics as
the paths of propagation of small disturbances. Let us suppose that the region
of the vwplane in which the function x(v, w) is not zero becomes a very narrow
strip along a characteristic. The derivatives of x in the direction transverse
to the characteristic then take a very wide range of values, since x diminishes
very rapidly in that direction. Such solutions x( v > w ) °f the equations of
motion must exist. For, regarded as a perturbation in the ^zuplane, they
satisfy the conditions of geometrical acoustics, and are therefore nonzero
along characteristics, as such perturbations must be.
It is clear from the foregoing that, for such a function x, the time t — dx[8w
will take an arbitrarily large range of values. The derivative of x along the
characteristic, however, is finite. Along a characteristic (for instance, a
characteristic V) we have
d/_ 1 dp da; 1 da?
dv pc dtv dv c dv
t It is most simply derived by using Cauchy's theorem in the theory of functions of a complex
variable. For an arbitrary function F(c{ u) we have
/ d \»i F(c + u) nm _j d \n~i F(c + u)
= 2»i[ —
cdc l c \ dc* I c
2mi J y/z{zc*y
where the integral is taken along a contour in the complex jarplane which encloses the point z — c 2 .
Putting now u = c\a and substituting in the integral ■s/z = 2£ — c, we obtain
1 (nl)l /^ + a)
2" 1 7mi J C n (Cc) n
where the contour of integration encloses the point £ = c; again applying Cauchy's theorem, we
have the result (98.10).
390 Onedimensional Gas Flow §98
The derivative of x with respect to v along a characteristic, which we denote
by —f( v )> i s therefore
^ = ^ + _^.^ = ^ + C ^L = _f( \
dv dv dw dv dv dw ~
Expressing the partial derivatives of x in terms of x and t by (98.1), we
obtain the relation x = (v + c)t+f(v) y i.e. the equation (94.5) for a simple
wave. The relation (94.4), which gives the relation between v and cina
simple wave, is necessarily satisfied, since/ is constant along a characteristic
r_.
We have shown in §97 that, if the solution of the equations of motion
reduces to constant flow in some part of the atfplane, then there must be a
simple wave in the adjoining regions. The motion described by the general
solution (98.8) must therefore be separated from a region of constant flow (in
particular, a region of gas at rest) by a simple wave. The boundary between
the simple wave and the general solution, like any boundary between two
analytically different solutions, is a characteristic. In solving particular
problems, the value of the function x (w> v) on this boundary characteristic
must be determined.
The "joining" condition at the boundary between the simple wave and the
general solution is obtained by substituting the expressions (98.1) for x and
t in the equation of the simple wave x — (v±c)t+f(v); this gives
— ± c^ +f(v) = 0.
dv dw
Moreover, in a simple wave (and therefore on the boundary characteristic),
we have dv = ± dpfpc = ± dwfc, or ± c = dw\dv. Substituting this in the
above condition, we obtain
8 X d x dw d x
— + T 7 +f(?) = — +f(v) = 0,
dv dw av dv
or, finally,
X =  J>)d*, (98.11)
which determines the required boundary value of x . In particular, if the
simple wave has a centre at the origin, i.e. if f(v) s 0, then x = constant;
since the function x is defined only to within an additive constant, we can
without loss of generality take x — on the boundary characteristic.
PROBLEMS
Problem 1. Determine the resulting flow when a centred rarefaction wave is reflected
from a solid wall.
Solution. Let the rarefaction wave be formed at the point x = at time t = 0, and
propagated in the positive ^direction ; it reaches the wall after a time t = l/c , where / is
§98
Arbitrary onedimensional gas flow
391
the distance to the wall. Fig. 75 shows the characteristics for the reflection of the wave.
In regions 1 and 1' the gas is at rest; in region 3 it moves with a constant velocity v = — C/.f
Region 2 is the incident rarefaction wave (with rectilinear characteristics C+), and region 5
is the reflected wave (with rectilinear characteristics C_). Region 4 is the "region of inter
action", in which the solution is required ; the linear characteristics become curved on entering
this region. The solution is entirely determined by the boundary conditions on the segments
ab and ac. On ab (i.e. on the wall) we must have v = for x = I; by (98.1), we hence obtain
the condition dx\dv = — / for v = 0. The boundary ac with the rarefaction wave is part of
a characteristic C_, and we therefore have c— $(y — l)v = c—v/(2n + l) = constant; since,
at the point a, v = and c — c , the constant is c . On this boundary x must be zero, so
that we have the condition x — for c— v/(2n + l) = c . It is easily seen that a function of
the form (98.9) which satisfies these conditions is
/(z«+iw d \ ni n I"/ V Y l n )
a)
and this gives the required solution.
Fig. 75
The equation of the characteristic ac is (see §96, Problem)
* = (2n+ l)c t + 2(n + l)/(fc //) (2n+1)/2(n+1) .
Its intersection with the characteristic Oc
xjt = c i{y+l)U = c 2(n+l)Ul{2n+l)
determines the time at which the incident wave disappears :
/(2»+l)»+l£ w
U. =
[(2n + iy C/]»+i'
In Fig. 75 it is assumed that U < 2c [(y+l); in the opposite case, the characteristic Oc
is in the negative ardirection (Fig. 76). The interaction of the incident and reflected waves
then lasts for an infinite time (not, as in Fig. 75, for a finite time).
f If the rarefaction wave is due to a piston which begins to move out of a'pipe at a constant velocity,
then U is the velocity of the piston.
392
Onedimensional Gas Flow
§99
The function (1) also describes the interaction between two equal centred rarefaction waves
which leave the points x = and x = 21 at time t = and are propagated towards each other ;
this is evident from symmetry (Fig. 77).
Fig. 76
Fig. 77
Problem 2. Derive the equation analogous to (98.3) for onedimensional isothermal flow
of a perfect gas.
Solution. For isothermal flow, the heat function w in Bernoulli's equation is replaced by
[X = J* dp/p = C T 2 j dp/p = ct 2 logp,
where ct z = {dp\dp)T is the square of the isothermal velocity of sound. For a perfect gas
ct = constant. Taking the quantity ju. (instead of w) as an independent variable, we obtain,
by the same method as in the text, the following linear equation with constant coefficients :
9 d2 X , d X d2 X n
d/j? dfj, dv 2
§99. The propagation of strong shock waves
Let us consider the propagation of a spherical shock wave of great intensity
resulting from a strong explosion, i.e. from the instantaneous release of a
§99 The propagation of strong shock waves 393
large quantity of energy (which we denote by E) in a small volume ; we sup
pose that the shock is propagated through a perfect gas (L. I. Sedov, 1946).
We shall consider the wave at relatively small distances from the source, so
that the amplitude is still large. These distances are, nevertheless, supposed
large in comparison with the dimensions of the source; this enables us to
assume that the energy E is generated at a single point (the origin).
If the shock wave is strong, the pressure discontinuity in it is very large.
We shall suppose that the pressure p% behind the discontinuity is so large,
compared with the pressure p\ of the undisturbed gas in front of it, that
PzlPi > (y+l)/(y— !)• This means that we can everywhere neglect pi in
comparison with /% and the density ratio pzjpi is equal to its limiting value
(y+l)/(yl);see§85.
Thus the gas flow pattern is entirely determined by two parameters: the
initial gas density pi, and the quantity of energy E generated in the explosion.
From these parameters and the two independent variables (the time t and
the radial coordinate r), we can form only one dimensionless combination,
which we write as
i = r( P1 /^)i/5. (99.1)
Consequently, we have a certain type of similarity flow.
We can say, first of all, that the position of the shock wave itself at every
instant must correspond to a certain constant value g o of the dimensionless
quantity £. This gives at once the manner in which the shock wave moves
with time; denoting by ro the distance of the shock from the origin, we have
ro = &(£* 2 //>i) 1/5 . (99.2)
From this we find the rate of propagation of the shock wave (relative to the
undisturbed gas, i.e. relative to a fixed coordinate system):
mi = drojdt = 2r jSt. (99.3)
It diminishes with time as t~*.
The gas pressure p2, the density p% and the velocity V2 = «i— «2 (relative
to a fixed coordinate system) just behind the discontinuity can be expressed
in terms of «i by means of the formulae derived in §85. According to (85.5)
and (85.6),f we have
v 2 = 2wi/(y+ 1), pz = pi(y + l)/(y 1), p 2 = 2p 1 u 1 *j(y+ 1). (99.4)
The density is constant in time, while V2 and p^ diminish as f~ s and t'*
respectively. We may also note that the pressure p2 due to the shock increases
with the total energy of the explosion as E*.
Let us next determine the gas flow throughout the region behind the shock.
t We here denote by aj and Ug the velocities of the shock wave, relative to the gas, given by for
mulae (85.6).
394 Onedimensional Gas Flow §99
Instead of the gas velocity v, the density p, and the pressure p, we introduce
dimensionless variables v\ p', p\ defined by
4 r y+1 8pi r 2
V, P = Pl Y —p', p= * ^ p'. (99.5)
5(y+l) * ' * * yV * 25(y+l) * 2
The quantities z/, p' andp' can be functions only of the dimensionless variable
£ . On the surface of discontinuity (i.e. for $ = ft) they must have the values
V ' = p' = p' = 1 for  = ft. (99.6)
The equations of centrallysymmetrical adiabatic gas flow are
dv 8v 1 dp dp d{pv) 2pv
1 v — = , — + + = 0,
dt or p dr dt dx r
(99.7)
Id 8 \ p
__ + *;_ log— = 0.
\dt 8r p?
The last equation is the equation of conservation of entropy, with the ex
pression (80.12) for the entropy of a perfect gas substituted. After substitut
ing (99.5), we obtain a set of ordinary differential equations for the functions
v\ p and p'. The integration of these equations is facilitated by the fact that
one integral can be obtained immediately, using the following arguments.
The fact that we have neglected the pressure pi of the undisturbed gas
means that we neglect the original energy of the gas in comparison with the
energy E which it acquires as a result of the explosion. It is therefore clear
that the total energy of the gas within the sphere bounded by the shock
wave is constant and equal to E. Furthermore, since we have a similarity
flow, it is evident that the energy of the gas inside any sphere of a smaller
radius, which increases with time in such a way that f = any constant (not
only ft), must remain constant; the radial velocity of points on this sphere
is v n = 2r/5* (cf. (99.3)).
It is easy to write down the equation which expresses the constancy of this
energy. On the one hand, an amount of energy d*. 4nr 2 pv(zo+%v 2 ) leaves
the sphere (whose area is 4nr 2 ) in time dt. On the other hand, the volume of
the sphere is increased in that time by dt . v n . 47rr 2 , and the energy of the gas
in this extra volume is d£ . 4rrr 2 pv n (e+%v 2 ). Equating the two expressions,
putting € = plp(y— 1) and w = ye, and introducing the dimensionless func
tions by (99.5), we obtain
p' (y+l2a>' 2
~ = ^ T' (99 ' 8 >
p Zyv — y— 1
which is the required integral. It automatically satisfies the boundary
condition (99.6) at the surface of discontinuity.
When the integral (99.8) is known, the integration of the equations is
§99 The propagation of strong shock waves 395
elementary though laborious. The second and third equations (99.7) give
dv' l y+l\dlog//
dlog£ \ 2 /dlog£
(99.9)
d / P'\ ^ 5(y + \)4v'
dlogi \° g p'r) 2v'(y+l)'
From these two equations we can express the derivatives d^'/d log £ and
d log p'/dv', by means of (99.8), as functions of v' only, and then an integra
tion with the boundary conditions (99.6) gives
\l) = * L TTy J [—y^\ •
V2yv'yl ]".[ 5(y+l)2(3yl)fl' l v *[ y+ 1 2v' y*
*  L^rJ L t^ J [~v=r~\ •
13y2_7 y +12 5(yl) 3
^1 = 77! 777^ — 77> v 2 = —  — — » »* =
(3yl)(2y+l)' 2y+l ' 2y+l'
13y 2 7y+12 1
"4 SB 7^ 77^ 7777; 77» "5
(2y)(3yl)(2y+l)' y2'
(99.10)
Formulae (99.8) and (99.10) give the complete solution of the problem. The
constant £o is determined by the condition
E ~]( ! r + £i) M '* r >
which states that the total energy of the gas is equal to the energy E of the
explosion. In terms of the dimensionless quantities, this condition becomes
Iq5 25( 3 21) J (^ V2 + ^W = 1. (99.11)
For instance, for air (y = f) this constant is £o = 1*033.
The ratios v\vi and pjpz as functions of rjr Q = f /£ are easily seen from the
above formulae to tend to zero as rjro > 0, in the manner
vfa ~ r/r , plp2 ~ (r/rofKvU; (99.12)
the ratio of pressures pjpz, however, tends to a constant, and the ratio of
temperatures therefore tends to infinity.
396
Onedimensional Gas Flow
§100
Fig. 78 shows the quantities vjv%, pjp2 and p/p2 as functions of r/ro for
air (y = 14). The very rapid decrease of the density into the sphere is
noticeable: almost all the gas is in a relatively thin layer behind the shock
wave. This is, of course, due to the fact that the gas on the surface of greatest
radius (ro) has a density six times the normal density, f
10
v/v 2 X J \
0'5
p/p 2 y
/__^y \
P/P2J
05
Fig. 78
10
§100. Shallowwater theory
There is a remarkable analogy between gas flow and the flow in a gravita
tional field of an incompressible fluid with a free surface, when the depth of
the fluid is small (compared with the characteristic dimensions of the problem,
such as the dimensions of the irregularities on the bottom of the vessel).
In this case the vertical component of the fluid velocity may be neglected
in comparison with its velocity parallel to the surface, and the latter may be
regarded as constant throughout the depth of the fluid. In this {hydraulic)
approximation, the fluid can be regarded as a "twodimensional" medium
having a definite velocity v at each point and also characterised at each point
by a quantity h, the depth of the fluid.
The corresponding general equations of motion differ from those obtained
in §13 only in that the changes in quantities during the motion need not be
assumed small, as they were in §13 in discussing long gravity waves of small
amplitude. Consequently, the secondorder velocity terms in Euler's equa
tion must be retained. In particular, for onedimensional flow in a channel,
f The results of calculations for other values of y are given by L. I. Sedov, Similarity and Dimen
sional Methods in Mechanics, Chapter IV, §11, CleaverHume Press, London 1959. The
corresponding problem with cylindrical symmetry is also discussed.
§100 Shallowwater theory 397
depending only on one coordinate x (and on the time), the equations are
8h 8{vh)
— + ±J = 0,
8t dX (100.1)
dv 8v 8h
1 v — = — g — ;
8t 8x 8x
the depth h is here assumed constant across the channel.
Long gravity waves are, in a general sense, small perturbations of the flow
now under consideration. The results of §13 show that such perturbations
are propagated relative to the fluid with a finite velocity, namely
C = y/(gh). (100.2)
This velocity here plays the part of the velocity of sound in gas dynamics.
Just as in §79, we can conclude that, if the fluid moves with velocities v < c
{streaming flow), the effect of the perturbations is propagated both upstream
and downstream. If the fluid moves with velocities v > c {shooting flow),
however, the effect of the perturbations is propagated only into certain regions
downstream.
The pressure/) (reckoned from the atmospheric pressure at the free surface)
varies with depth in the fluid according to the hydrostatic law/) = pg{h—z),
where z is the height above the bottom. It is useful to note that, if we intro
duce the quantities
h
p = P h, p = jpdz = \pgh* = gpfr, (100.3)
o
then equations (100.1) become
85 8{vp) 8v dv 1 8b
£ + ^ =0, — + v— =   ■/, (100.4)
8t 8x 8t 8x p 8x
which are formally identical with the equations of adiabatic flow of a perfect
gas with y = 2 {p ~ p~ 2 ). This enables us to apply immediately to shallow
water theory all the results of gas dynamics for flow in the absence of shock
waves. If shock waves are present, however, the results of shallowwater
theory differ from those of perfectgas dynamics.
A "shock wave" in a fluid in a channel is a discontinuity in the fluid height
h s and therefore in the fluid velocity v (what is called a hydraulic jump).
The relations between the values of the quantities on the two sides of the
discontinuity can be obtained from the conditions of continuity of the fluxes
of mass and momentum. The mass flux density (per unit width of the
channel) is/ = pvh. The momentum flux density is obtained by integrating
p f pv* over the depth of the channel, and is
f (p+pv 2 )dz = \pgh 2 + pv 2 h.
398 Onedimensional Gas Flow §100
The conditions of continuity therefore give two equations:
vifa = v 2 h 2 , (100.5)
»i 2 Ai+ W = v 2 2 h 2 + ±gh 2 2 . (100.6)
These give the relations between the four quantities vi, v 2 , hi, h 2 , two of which
can be specified arbitrarily. Expressing the velocities vi and v 2 in terms of
the heights h\ and h 2 , we obtain
*>i 2 = &h 2 (hi+h 2 )lhi, v 2 2 = ighi(hi + h 2 )/h 2 . (100.7)
The energy fluxes on the two sides of the discontinuity are not the same, and
their difference is the amount of energy dissipated in the discontinuity per
unit time. The energy flux density in the channel is
n
q = J (  + frApvdz = ij(gh+v 2 ).
Using (100.7), we find the difference to be
qiq 2 = gj(hi 2 + h 2 2 )(h 2 hi)/4hih 2 .
Let the fluid move through the discontinuity from side 1 to side 2. Then the
fact that energy is dissipated means that qi — q 2 > 0, and we conclude that
h 2 > h u (100.8)
i.e. the fluid moves from the smaller to the greater height. We then can
deduce from (100.7) that
vi > ci = V(gh), v 2 < c 2 = VW, (100.9)
in complete analogy to the results for shock waves in gas dynamics. The
inequalities (100.9) could also be derived as the necessary conditions for the
discontinuity to be stable, as in §84.
CHAPTER XI
THE INTERSECTION OF
SURFACES OF DISCONTINUITY
§101. Rarefaction waves
The line of intersection of two shock waves is, mathematically, a singular
line of two functions describing the gas flow. The vertex of an acute angle on
the surface of a body past which the gas flows is always such a singular line.
It is found that the gas flow near the singular line can be investigated in a
general manner (L. Prandtl and T. Meyer, 1908).
In considering the region near a small segment of the singular line, we may
regard the latter as a straight line, which we take as the #axis in a system of
cylindrical coordinates r, </», z. Near the singular line, all quantities depend
considerably on the angle </>, but their dependence on the coordinate r
is only slight, and for sufficiently small r it can be neglected. The dependence
on the coordinate z is also unimportant; the change in the flow pattern over
a small segment of the singular line may be neglected.
Thus we have to investigate a steady flow in which all quantities are func
tions of <f> only. The equation of conservation of entropy, v«grad$ = 0,
gives v$ dy/d<£ = 0, whence $ = constant,f i.e. the flow is isentropic. In
Euler's equation we can therefore replace grad pjp by grad to: (vgrad)v
= — grad zv. In cylindrical coordinates, we have three equations :
^^V_V_ VfdVf VfVj, l^da; dv z _
r d<f> r ' r d<f> r r d<f>' d<f>
From the last of these we have v z = constant, and without loss of generality
we can put v z = 0, regarding the flow as twodimensional; this is simply a
matter of suitably defining the velocity of the coordinate system along the
2axis. The first two equations can be written
^ = d«v/<ty, (101.1)
dv$ \ \ dp dw
d(f> / p d<f> d<f>
Substituting (101.1) in (101.2), we have
dvf dv r dw
V,f d4 + Vr d$ = ~ d0 '
t If »* = 0, we easily deduce from the equations of motion given below that v r — 0, v t ^ 0. Such
a flow would correspond to the intersection of surfaces of tangential discontinuity (with a discontinuous
velocity v t ), and is of no interest, since such discontinuities are unstable.
399
400 The Intersection of Surfaces of Discontinuity §101
or, integrating,
w + hiv^+Vr 2 ) = constant. (101.3)
We may notice that equation (101.1) implies that curl v = 0, i.e. we have
potential flow, as a result of which Bernoulli's equation (101.3) holds.
Next, the equation of continuity, div(pv) = 0, gives
Using (101.2), we obtain
(?)('<)*
/4 A
Fig. 79
The derivative dp [dp, or more correctly (dp/d/>) s , is just the square of the
velocity of sound. Thus
(IHf 1 ?) ^
This equation can be satisfied in either of two ways. Firstly, we may have
dv^jdcji+Vr = 0. Then, from (101.2), p — constant and p = constant, and
from (101.3) we find that v 2 = v r 2 + v^ 2 = constant, i.e. the velocity is constant
in magnitude. It is easy to see that in this case the velocity is constant in
direction also. The angle x between the velocity and some given direction in
the plane of the motion is (Fig. 79)
x = <£ + tani(V^) ( 10L6 )
Differentiating this expression with respect to <j> and using formulae (101.1)
and (101.2), we easily obtain
d x /d«£ = (vrlpv^dpldc/,. (101.7)
Since p = constant, it follows that x = constant. Thus, if the first factor
in (101.5) is zero, we have the trivial solution of a uniform flow.
* f ^ . (101.9)
§101 Rarefaction waves 401
Secondly, equation (101.5) can be satisfied by putting X — v^jc* = 0,
i.e. Vf = ±c. The radial velocity is given by (101.3). Denoting the constant
in that equation by wq, we find that
«V = ±c, v r = ±\Z[2(w w)c 2 ].
In this solution, the velocity component v^ perpendicular to the radius
vector is equal to the local velocity of sound at every point. The total velocity
v = V(*V 2 + ^r 2 ) therefore exceeds that of sound. Both the magnitude and
the direction of the velocity are different at different points. Since the velocity
of sound cannot vanish, it is clear that the function v^,(<f>), which is continuous,
must everywhere be + c, or else everywhere — c. By measuring the angle <f>
in the appropriate direction, we can take v^ = c. We shall see below that the
choice of the sign of v r follows from physical considerations, and that the
plus sign must be taken. Thus
«V = c, v r = V[2(mw)c*]. (101.8)
From the equation of continuity (101.4) we have d<f> = — d(pv^)fpv r . Sub
stituting (101.8) and integrating, we have
d(pc)
p\/[2(zvow)c 2 ]
If the equation of state of the gas and the adiabatic equation are known (we
recall that s is constant), this formula can be used to determine all quantities
as functions of the angle <f>. Thus formulae (101.8) and (101.9) completely
determine the gas flow.
Let us now study in more detail the solution which we have obtained.
First of all, we notice that the straight lines <f> = constant intersect the stream
lines at every point at the Mach angle (whose sine is v^jv = cfv), i.e. they
are characteristics. Thus one family of characteristics (in the ryplane)
is a pencil of straight lines through the singular point, and has an important
property in this case: all quantities are constant along each characteristic.
In this respect the solution concerned plays the same part in the theory of
steady twodimensional flow as does the similarity flow discussed in §92
in the theory of nonsteady onedimensional flow. We shall return to this
point in §107.
It is seen from (101.9) that (pc)' < 0, the prime denoting differentiation
with respect to <f>. Putting (pc)' = pd(pc)jdp and noticing that the derivative
d(pc)jdp is positive (see (92.9)), we find that p < 0, and therefore so are the
derivatives p' = c 2 p and zv' = p'jp. Next, from the fact that w' is negative
it follows that the velocity v = ■\/\2{v)q — w)] increases with <f>. Finally,
from (101.7), x > 0. Thus we have
dp/d<p < 0, dp/dcf> < 0, dv/dcf> > 0, d x /d<£ > 0. (101.10)
In other words, when we go round the singular point in the direction of flow,
the density and pressure decrease, while the magnitude of the velocity in
creases and its direction rotates in the direction of flow.
402 The Intersection of Surfaces of Discontinuity §101
The flow just described is often called a rarefaction wave, and we shall
use this name in what follows.
It is easy to see that a rarefaction wave cannot exist throughout the region
surrounding the singular point. For, since v increases monotonically with <f>,
a complete circuit round the origin (i.e. a change of <j> by 2tt) would give a
value for v different from the initial one, which is impossible. For this reason,
the actual pattern of flow round the singular line must be composed of a
series of sectors separated by planes <j> = constant which are surfaces of
discontinuity. In each of these regions we have either a rarefaction wave or a
flow with constant velocity. The number and nature of these regions for
various particular cases will be established in the following sections . Here we
shall simply mention that the boundary between a rarefaction wave and a
uniform flow must be a weak discontinuity: it cannot be a tangential
discontinuity (of v r ), since the normal velocity component v^ = c does not
vanish on it. Nor can it be a shock wave, since the normal velocity component
v^ must be greater than the velocity of sound on one side of such a dis
continuity and smaller on the other side, whereas in our problem we always
have Vq = c on one side of the boundary.
An important conclusion can be drawn from the foregoing. Disturbances
which cause weak discontinuities evidently leave the singular line (the #axis)
and are propagated away from it. This means that the weak discontinuities
bounding the rarefaction wave must be ones which leave this line, i.e. the
velocity component v r tangential to the weak discontinuity must be positive.
This justifies the choice of the sign of v r made in (101.8).
Let us now apply these formulae to a perfect gas. In such a gas
w ~ c2 l(v~ 1)> while the equation of the Poisson adiabatic can be written
p C 2/(yD = constant, ^ c 2y/(ri)  constant; (101.11)
cf. (92.13). Using these formulae, we can put the integral (101.9) in the form
fy+1 r dc
—JS!
y\ J V(c* 2 c 2 )'
where c# is the critical velocity (see (80.14)). Hence
rr>«s~ J —
<f> = / cos 1 h constant,
Vy1 c*
or, if we measure <j> in such a way that the constant is zero,
^ = c = c m cos V[(y l)/(y + l)tf. (101.12)
According to formula (101.8) we therefore have
^ = y^r* sin V^ < 10U3 >
Next, using the Poisson adiabatic equation in the form (101.11), we can find
§101 Rarefaction waves
the pressure as a function of the angle <f> :
(Yl
p = p m cogyHr
i) /^.
Vy+r
Finally, we have for the angle x (101.6)
1 /yl
cot /
y+1 Vy +
the angles x and <f> being measured from the same initial line.
403
(101.14)
*  ^^(yyw^)' (i ° u5)
09
08
07
^' 06
°1 05
^ 0*4
140°
120°
100°
80° a
scjv
p/p
03
02
01
sx
oO
40°
20
4C
f
8(
F
12
4
IG. i
iO°
so
«
50°
2C
X>°
Since we must have v r > 0, c> 0, the angle <f> in these formulae can vary
only between and <^nax, where
<?W = W[(y+l)/(yl)]
(101.16)
This means that the rarefaction wave can occupy a sector whose angle does
not exceed <£max; for a diatomic gas (air, for example), this angle is 2193°.
When <j> varies from to <£max, the angle x varies from \n to <£ma X . Thus the
direction of the velocity in the rarefaction wave can turn through an angle
not exceeding <£maxih (= 1293° for air).
For <f> = ^max the pressure is zero. In other words, if the rarefaction wave
occupies the maximum angle, the weak discontinuity on one side is a boundary
with a vacuum, and is, of course, a streamline; we have v$ = c = 0,
v r  v = V[(y+l)/(yl)] c # = ^max, i.e. the velocity is radial and attains
its limiting value © max (see §80).
Fig. 80 shows graphs of pip*, cjv and x as functions of the angle <f> for
air (y = 14).
404
The Intersection of Surfaces of Discontinuity
§101
It is useful to note the form of the curve in the ©a^plane defined by
formulae (101.12) and (101.13) (called the velocity hodograph). It is an arc of
an epicycloid between circles of radii v = c # and v = c%\Z[(y+l)l(y— 1)]
= *>max (Fig. 81).
Fig. 81
PROBLEMS
Problem 1 . Determine the form of the streamlines in a rarefaction wave.
Solution. The equation of the streamlines for twodimensional flow is, in polar co
ordinates, dr/vr = rd<£/w<£. Substituting (101.12) and (101.13) and integrating, we obtain
r = r cos(r+i)/(yi)^/[(yl)/( y+ l)]^
These streamlines form a family of similar curves concave toward the origin, which is the
centre of similarity.
Problem 2. Determine the maximum possible angle between the weak discontinuities
bounding a rarefaction wave, for given values v u c x of the gas velocity and the velocity of
sound at one discontinuity.
Solution. The angle <j> corresponding to the first discontinuity is, by (101.12),
<f>l =
y+1
C\
COS"
y1 c*
The value of <f> 2 is <f>m&x, so that the angle required is
y+1
fc* 1 = Jhri
sin"
ifl
The critical velocity c* is given in terms of v x and c x by Bernoulli's equation :
ZOi + ivi 2 =
c?
> 1
+ w =
7+1
2(71)'
The maximum possible angle through which the gas velocity can turn in a rarefaction wave
is accordingly, by (101.15), the difference Xmax = x(«£i) — xifa)'
Xmax
y+1 . , C\ . Ci
— sin 1 sin 1 — .
y— 1 c# vi
§102 The intersection of shock waves 405
As a function of vJc x , Xmax is greatest for v^Cx = 1 :
(y+1
y\
i i i y+1 1
Xmax = fw /   1
For v t /ci > oo, Xmax tends to zero:
Xmax
2 Ci
y— 1 V\
§102. The intersection of shock waves
Shock waves can intersect along a line. In considering the flow near a small
segment of this line, we can assume that it is a straight line, and that the sur
faces of discontinuity are planes. It is therefore sufficient to discuss the inter
section of plane shock waves.
The line of intersection of two discontinuities is, mathematically, a singular
line, as has already been mentioned at the beginning of §101. The flow
pattern near this line consists of a number of sectors, in each of which we have
either uniform flow or a rarefaction wave of the kind described in §101.
It is possible to give a general classification of the possible types of intersec
tion of surfaces of discontinuity (L. Landau 1944).
First of all, we must make the following remark. If the gas flow on both
sides of a shock wave is supersonic, then (as mentioned at the beginning of
§86) we can speak of the "direction" of the shock wave, and accordingly
distinguish shock waves leaving the line of intersection from those reaching it.
In the former case, the tangential velocity component is directed away from
the line of intersection, and we can say that the disturbances which cause the
discontinuity leave this line. In the latter case, the perturbations leave a
point not on the line of intersection.
If the flow on one side of the shock wave is subsonic, then disturbances are
propagated in both directions along its surface, and the "direction" of the
shock has, strictly, no meaning. In the arguments given below, however,
what is important is that disturbances leaving the point of intersection can be
propagated along such a discontinuity. In this sense, such shock waves play
the same part in the following discussion as the purely supersonic shocks
which leave the intersection, and we shall include both kinds in the term
"shocks which leave the intersection".
Figs. 8286 show the flow patterns in a plane perpendicular to the line of
intersection. We can assume, without loss of generality, that the flow occurs
in this plane. The velocity component parallel to the line of intersection
(which lies in all the planes of discontinuity) must be the same in all regions
round the line of intersection, and can therefore be made to vanish by an
appropriate choice of the coordinate system.
It is easy to see that there can be no intersection of shock waves in which no
shock reaches the intersection. For instance, in the intersection of two shock
waves leaving the intersection, shown in Fig. 82a, the streamlines of the flow
14
406 The Intersection of Surfaces of Discontinuity §102
incident from the left would deviate in opposite directions, whereas the
velocity should be constant throughout region 2, and this difficulty cannot be
overcome by adding any further discontinuities in region 2.f Similarly, we
can see that the intersection of a shock wave and a rarefaction wave both
leaving the intersection, shown in Fig. 82b, is impossible; although the
velocity in region 2 can be constant in direction, the pressure cannot be con
stant, since it increases in a shock wave but decreases in a rarefaction wave.
Shock Weak Tangential Streamline
wave discontinuity discontinuity
Fig. 82
Next, since the intersection cannot affect shock waves reaching it, the
simultaneous intersection (along a common line) of more than two such waves,
which are due to other causes, would be an improbable coincidence. Thus
only one or two shock waves can reach the intersection.
The following fact is very important. The gas flowing past a point of
intersection can pass through only one shock or rarefaction wave leaving this
point. For example, let the gas pass through two successive shock waves
leaving the point O, as shown in Fig. 82c. Since the normal velocity com
ponent V2n behind the shock Oa is less than C2, the velocity component in
region 2 normal to the shock Ob must also be less than c%, in contradiction to
a fundamental property of shock waves. Similarly, we can see that the gas
cannot pass through two successive rarefaction waves, or a shock wave and a
rarefaction wave, leaving the point O.
These arguments evidently cannot be extended to shock waves reaching the
point of intersection.
We can now proceed to enumerate the possible types of intersection. Fig.
83 shows an intersection involving one shock wave Oa reaching it and two
shock waves Ob, Oc leaving it. This case may be regarded as the splitting of
one shock wave into two. J It is easy to see that, besides the two shock waves
t In order not to encumber the discussion with repetitive arguments, we shall not give similar
considerations for cases where there are regions of subsonic flow and the shock leaving the inter
section is actually a shock wave bounded by a subsonic region.
% It should be noticed that a shock wave cannot divide into a shock and a rarefaction wave; it is
easily seen that the changes in the pressure and the direction of the velocities in the two waves leaving
cannot be reconciled.
§102 The intersection of shock waves 407
leaving, there must be formed a tangential discontinuity Od lying between
them, which separates the gas flowing through Ob from that flowing through
Oc.f For the shock Oa is due to other causes, and is therefore completely
defined. This means that the thermodynamic quantities (p and p, say) and
the velocity v have given values in regions 1 and 2. There remain at our dis
posal, therefore, only two quantities (the angles giving the directions of the
discontinuities Ob and Oc) with which to satisfy, in general, four conditions
(the constancy of p, p and two velocity components) in the region 34,
which would have to be satisfied in the absence of the tangential discon
tinuity Od. The addition of the latter, however, reduces the number of
conditions to two (the constancy of the pressure and of the direction of the
velocity).
Fig. 83
An arbitrary shock wave, however, cannot divide in this manner. A shock
wave reaching the intersection is defined by two parameters (for a given
thermodynamic state of gas 1), say the Mach number Mi of the incident
stream and the ratio of pressures pi[p2 It can divide in two only in a certain
region in the plane of these two parameters.^
Intersections involving two shock waves reaching them can be regarded as
"collisions" of two shocks due to other causes. Here two essentially different
cases are possible, as shown in Fig. 84.
In the first case, the collision of two shock waves results in two other
shock waves leaving the point of intersection. If all the necessary conditions
t As usual, the tangential discontinuity in reality becomes a turbulent region.
% The determination of this region involves very laborious algebraic calculations. The results that
have been published (see, for example, R. Courant and K. O. Friedrichs, Supersonic Flow and
Shock Waves, Interscience, New York 1948), unfortunately, are largely invalidated by the fact that
they make no distinction between shock waves reaching and leaving the intersection. The ternary
configurations therefore include also those where two shock waves reach the intersection and one
leaves it. This, however, is the intersection of two shocks due to other causes, and therefore reaching
the point of intersection with given values of all parameters. Their "fusion" into one shock is possible
only when these arbitrary parameters are related in a certain way, and this would be an improbable
coincidence.
408
The Intersection of Surfaces of Discontinuity
§102
are to be fulfilled, a tangential discontinuity must again be formed, and it
must lie between the two resulting shock waves.
In the second case, instead of two shock waves, there are formed one shock
wave and one rarefaction wave.
Fig. 84
Two colliding shock waves are defined by three parameters (for instance,
Mi and the ratios p\jp2, Pijpz) The types of intersection just described are
possible only for certain ranges of values of these parameters. If the values
of the parameters do not lie in these regions, the collision of the shock waves
must be preceded by their breaking up.
Fig. 85 shows the reflection of a shock wave from the boundary between
gas in motion and gas at rest. Region 5 contains gas at rest, separated from
the gas in motion by a tangential discontinuity. In the two regions 1 and 4
adjoining it, the pressure must be the same and equal top$. Since the pressure
increases in a shock wave, it is clear that the shock wave must be reflected
from the tangential discontinuity as a rarefaction wave 3, which reduces the
pressure to its initial value.
Finally, we may briefly discuss the intersection of a shock wave with a weak
discontinuity arriving from an external source. Here two cases can occur,
§102
The intersection of shock waves
409
according as the flow behind the shock wave is supersonic or subsonic. In
the former case (Fig. 86a), the weak discontinuity is "refracted" at the shock
wave into the space behind the latter; the shock itself is not refracted at the
intersection, but has a singularity of a higher order, like that at a weak dis
continuity. Moreover, the entropy change in the shock wave must cause
behind it a "weak tangential discontinuity", at which the derivatives of the
entropy are discontinuous.
Fig. 85
Weak
tangential
discontinuity
Weak
discontinuity
Fig. 86
If, however, the flow becomes subsonic behind the shock wave, the weak
discontinuity cannot penetrate into this region, and it ceases at the point of
intersection (Fig. 86b). The latter is now a singular point; it can be shown
that the velocity distribution behind the shock wave has a logarithmic sin
gularity at this point. Furthermore, as in the previous case, a weak tangential
discontinuity of the entropy must occur behind the shock wave.f
t A detailed qualitative and quantitative analysis of the possible types of intersection of shock waves
with weak discontinuities is given by S .P. D'yakov, Zhurnal experimental' noli teoreticheskoi fiziki 33
948, 962, 1957; Soviet Physics JETP 6 (33), 729, 739, 1958; Doklady Akademii Nauk SSSR 99
921, 1954.
410 The Intersection of Surfaces of Discontinuity §103
§103. The intersection of shock waves with a solid surface
An important part in the phenomenon of steady intersection of shock waves
with the surface of a body is played by their interaction with the boundary
layer. This interaction is very complex, and has not yet been sufficiently
investigated, either experimentally or theoretically. However, simple general
arguments enable us to obtain some important results, which we shall now
expound, f
The pressure is discontinuous in a shock wave, and increases in the direc
tion of motion of the gas. Hence, if the shock wave intersects the surface,
there must be a finite increment of pressure over a very short distance near
the place of intersection, i.e. there must be a very large positive pressure
gradient. We know, however, that such a rapid increase in pressure cannot
occur near a solid wall (see the end of §40) ; it would cause separation, and the
pattern of flow round the body is changed in such a way that the shock wave
moves away to a sufficient distance from the surface.
These arguments, however, do not apply when the shock wave is weak.
It is clear from the proof given at the end of §40 that the impossibility of
a positive pressure discontinuity at the boundary layer is a consequence of
the assumption that this discontinuity is large: it must exceed a certain
limit depending on the value of R, which diminishes when R increases. J
Thus we reach the following important conclusions. The steady inter
section of strong shock waves with a solid surface is impossible. A solid
surface can intersect only weak shock waves, and the limiting intensity is the
smaller, the greater R. The maximum permissible intensity of the shock wave
also depends on whether the boundary layer is laminar or turbulent. If the
boundary layer is turbulent, the onset of separation is retarded (§45). In a
turbulent boundary layer, therefore, stronger shock waves can leave the
surface of the body than in a laminar boundary layer. ff
To avoid misunderstanding, it should be emphasised that these arguments
rely on the fact that the boundary layer exists in front of the shock wave
(i.e. upstream of it). The results obtained therefore relate, in particular, to
shock waves which leave the trailing edge, but not to those which leave the
leading edge, of the body; the latter can occur, for instance in flow past an
acuteangled wedge, a case which is discussed in detail in §104. In the latter
case the gas reaches the vertex of the angle from outside, i.e. from a region
in which there is no boundary layer. It is therefore clear that the present
f The boundary layer necessarily contains a subsonic part adjoining the surface, into which the
shock wave cannot penetrate. In speaking of the intersection, we ignore this fact, which does not affect
the following discussion.
% In §40, Problem, we have determined the smallest pressure change Ap over a distance Ax which
can cause separation in a laminar boundary layer. In the present application, we are concerned with
the pressure change over a distance of the order of the thickness 8 of the boundary layer, and obtain
the following law governing the decrease of Ap when the Reynolds number increases :
*PlP ~ l/R** ~ l/R# f .
ft The existing published data do not enable us to specify the maximum permissible intensity.
§103 The intersection of shock waves with a solid surface 411
arguments do not deny that shock waves can occur which leave the vertex of
such an angle.
In subsonic flow, separation can occur only when the pressure in the main
stream increases downstream along the surface. In supersonic flow, however,
it is found that separation can occur even when the pressure decreases down
stream. Such a phenomenon can occur by the combination of a weak shock
wave with a separation, the pressure increase necessary for separation taking
place in the shock wave ; the pressure may either increase or decrease down
stream in the region in front of the shock wave.
Fig. 87
The data at present existing do not enable us to give a detailed picture of the
complex phenomena involved in the "reflection" of a shock wave from the
subsonic part of a boundary layer (or from the turbulent region beyond the
line of separation). An important part in these phenomena must be played
by the fact that the disturbances due to the shock wave can be propagated
both upstream and downstream through the subsonic part of the boundary
layer, and can cause further discontinuities in it. In particular, the formation
of another weak shock wave upstream may result in separation, which "dis
places" a strong shock wave incident on the surface from outside. In Fig.
87, the line a is the incident shock wave, and b the shock wave formed up
stream, which causes separation at the point O. When the incident shock is
"reflected" from the subsonic part of the turbulent region, we should expect,
in particular, that a rarefaction wave would be formed.
All the above discussion relates only to a steady intersection, with the shock
wave and the body at relative rest. Let us now consider nonsteady intersec
tions, when a moving shock wave is incident on a solid body, so that the line
of intersection moves on the surface. Such an intersection is accompanied by
reflection of the shock wave: besides the incident wave, a reflected wave
leaving the body is formed.
We shall examine the phenomenon in a system of coordinates which moves
with the line of intersection; in this system the shock waves are steady.
The simplest type of reflection occurs when the reflected wave leaves the line
of intersection itself; this is called regular reflection (Fig. 88). If the angle of
incidence ai and the intensity of the incident shock are given, the flow in
region 2 is uniquely determined. The gas velocity in the reflected shock must
412 The Intersection of Surfaces of Discontinuity §103
be turned through an angle such that it is again parallel to the surface. When
this angle is given, the position and intensity of the reflected shock are ob
tained from the equation of the shock polar. For a given angle, the shock
polar determines two different shock waves, those of the weak and strong
families (§86). Experimental results show that in fact the reflected shock
always belongs to the weak family, and we shall assume this in what follows.
It should be pointed out that, when the intensity of the incident shock tends
to zero, the intensity of the reflected shock v then tends to zero also, and the
angle of reflection <X2 tends to the angle of incidence oci, as we should expect in
accordance with the acoustic approximation. In the limit ai > 0, the
reflected shock of the weak family passes continuously into the shock ob
tained when a shock wave is incident "frontally" (§93. Problem 1).
3
y77777777777J77777777777777777777.
Fig. 88
The mathematical calculations for regular reflection (in a perfect gas) offer
no difficulty in principle, but the algebra is extremely laborious. Here we
shall give only some of the results.^
It is clear from the general properties of the shock polar that regular
reflection is not possible for arbitrary values of the parameters of the incident
wave (the angle of incidence ai and the ratio pi\p\). For a given ratio P2JP1
there is a maximum possible angle ai&,J and for ai > <xifc regular reflection
is impossible. As p^jpi > oo, the maximum angle tends to sin 1 (l/y)
(= 40° for air). As p2Jpi »■ 1, «i& tends to 90°, i.e. regular reflection is
possible for any angle of incidence. Fig. 89 shows ai^ as a function of Pijpz
for air.
The angle of reflection <*2 is not in general the same as the angle of incidence.
There is a value a* of the angle of incidence such that, if ai < a # ,the angle
of reflection <X2 < ai; if ai> a*, on the other hand, <*2 > ai. The value of
a* is % cos 1 l(y— 1) (= 392° for air); it does not depend on the intensity
of the incident wave.
t A more detailed account of the reflection of shock waves is given by R. Courant and K. O.
Friedrichs, Supersonic Flow and Shock Waves, Interscience, New York 1948, and by W. Bleakney
and A. H. Taub, Reviews of Modern Physics 21, 584, 1949.
The solution of complex problems concerning the regular reflection of a shock wave at almost nor
mal incidence on the vertex of an angle close to 180°, and the diffraction of a shock wave at glancing
incidence on the vertex of a similar angle, has been given by M. J. Lighthill (Proceedings of the
Royal Society A198, 454, 1949; 200, 554, 1950).
J This is the value of the angle of incidence for which the strong and weak reflected shocks
coincide.
§104
Supersonic flow round an angle
413
For ai > auk regular reflection is impossible, and the incident shock wave
must break up at a distance from the surface, so that we have the pattern shown
in Fig. 90, with three shock waves, and a tangential discontinuity leaving the
point where the incident shock wave divides.
90"
80°
70°
60°
* 50°
40°
30°
20°
10°
10 09 08 07 06 05 04 03 02 01
P\/P2
Fi
G. S
9
a»
77777777777777777777777
Fig. 90
§104. Supersonic flow round an angle
In investigating the flow near the vertex of an angle on the surface, it is
again sufficient to consider small portions of the vertex and suppose it
straight, the angle being formed by two intersecting planes. We shall speak
of flow outside an angle if the angle is greater than 7r, and of flow inside an
angle if it is less than it.
Subsonic flow past an angle is not essentially different from the flow of an
incompressible fluid. Supersonic flow, however, is entirely different; an
important property of it is the occurrence of discontinuities leaving the vertex
of the angle.
Let us first consider the possible flow patterns when a supersonic gas
stream reaches the vertex along one of the sides of the angle. In accordance
with the general properties of supersonic flow, the stream remains uniform
up to the vertex. The turning of the stream into the direction parallel to
414 The Intersection of Surfaces of Discontinuity §104
the other side of the angle occurs in a rarefaction wave leaving the vertex,
and the flow pattern consists of three regions separated by weak discon
tinuities {Oa and Ob in Fig. 91): the uniform gas stream 1 moving along
the side AO is turned into the rarefaction wave 2 and then moves, again
with constant velocity, along the other side of the angle. It should be
noticed that no turbulent region is formed; in a similar flow of an incom
pressible fluid, on the other hand, a turbulent region must be formed, with
a line of separation at the vertex of the angle (Fig. 16, §35).
0/>
7777777777777777777,
(a)
0/
^77777777777777777^^
(b)
Let v\ be the velocity of the incident stream (1 in Fig. 91), and c\ the
velocity of sound in it. The position of the weak discontinuity Oa is deter
mined immediately from the Mach number Mi = v\\c\ by the condition
that it intersects the streamlines at the Mach angle. The changes in velocity
and pressure in the rarefaction wave are determined by formulae (101.12)
(101.15); all that is needed is the direction from which the angle </> in these
formulae is to be measured. The straight line <j> = corresponds to
v = c = c*; for Mi > 1, there is in fact no such line, since vjc > 1 every
where. However, if the rarefaction wave is imagined to be formally extended
into the region to the left of Oa, we can use formula (101.12), and we find
§104 Supersonic flow round an angle 415
that the discontinuity Oa must correspond to a value of <f> given by
ly+1
A A +1 l Cl
<f>l = / rcos i— ,
Vy1 c*
and that <f> must increase from Oa to 03. The position of the discontinuity Ob
is determined by the fact that the direction of the velocity becomes parallel to
the side OB of the angle.
/ 2
1
777777777777777777.
1
777777777777777777,
Fig. 92
Fig. 93
The angle through which the stream turns in the rarefaction wave cannot
exceed the value ^max determined in §101, Problem 2. If the angle /? round
which the flow occurs is less than it— xmax> the rarefaction wave cannot turn
the stream through the necessary angle, and we have the flow pattern shown
in Fig. 91b. The rarefaction in the wave 2 then proceeds to zero pressure
(reached on the line Ob\ so that the rarefaction wave is separated from the
wall by a vacuum (region 3).
The flow pattern described above is not the only possible one, however.
Figs. 92 and 93 show patterns in which a region of gas at rest adjoins the
second side of the angle, this region being separated from the moving gas by
a tangential discontinuity; as usual, this becomes a turbulent region, so that
the case considered corresponds to the presence of separation.f The stream
is turned through a certain angle in a rarefaction wave (Fig. 92) or in a shock
wave (Fig. 93). The latter case, however, is possible only if the shock wave is
not too strong (in accordance with the general considerations given in §103).
Which of these flow patterns will occur in any particular case depends in
general on the conditions far from the angle. For instance, when gas flows
out of a nozzle (the vertex of the angle being here the edge of the outlet), the
relation between the pressure p\ of the outgoing gas and the pressure p e of
the external medium is of importance. If p e < pi, the flow is of the type shown
in Fig. 92; the position and angle of the rarefaction wave are then determined
f According to experimental results, the compressibility of the gas somewhat diminishes the
angle of the turbulent region resulting from the tangential discontinuity.
416 The Intersection of Surf aces of Discontinuity §104
by the condition that the pressure in regions 3 and 4 is equal to p e . The
smaller p e , the greater the angle through which the stream must be turned.
If, however, the angle /? (Fig. 92) is large, the gas pressure cannot reach the
required value p e ; the direction of the velocity becomes parallel to the side
OB of the angle before the pressure falls to p e . The flow near the edge of the
outlet will then be as shown in Fig. 90. The pressure near the outer side
OB of the outlet is entirely determined by the angle /5, and does not depend
on the pressure p e \ the final decrease of the pressure to p e occurs only at a
distance from the outlet.
77777777777777' "
Fig. 94
If p e > pi, on the other hand, the flow round the edge of the outlet is of
the type shown in Fig. 93, with a shock wave which leaves the edge and raises
the pressure from^i to p e . This is possible, however, only if the difference
between p e and p\ is not too large, i.e. the shock wave is not too strong;
otherwise there is separation at the inner surface of the nozzle, and the shock
wave moves into the nozzle, in the manner described in §90.
Next, let us consider flow inside an angle. In the subsonic case such a flow
is accompanied by separation at a point ahead of the vertex (see the end of
§40). For a supersonic incident flow, however, the change in direction may
be effected by a shock wave leaving the vertex (Fig. 94). Here it must again
be mentioned that such a simple separationless flow pattern is possible only
if the shock wave is not too strong. Its intensity increases with the angle %
through which the stream is turned, and we can therefore say that separation
less flow is possible only when % is not too large.
Let us now consider the flow pattern which results when a free supersonic
stream is incident on the vertex of an angle (Fig. 95). The stream is turned
into directions parallel to the sides of the angle by shock waves leaving the
vertex. As has been shown in §103, this is the exceptional case where a shock
wave of arbitrary intensity can leave a solid surface.
If we know the velocities vi and c\ in the incident stream, we can deter
mine the positions of the shock waves and the gas flow in the regions behind
them. The direction of the velocity V2 must be parallel to the side OA of the
angle : V2yfv2 X = tan x Thus V2 and the angle <f> giving the position of the
shock wave can be determined immediately from the shock polar, using a
chord through the origin at the known angle x to the axis of abscissae (Fig.
50), as explained in §86. We have seen that, for a given x> the shock polar
gives two different shock waves, with different values of <f>. One of these
§104 Supersonic flow round an angle 4YJ
(corresponding to the point B in Fig. 50) is the weaker, and in general leaves
the flow supersonic; the other, stronger, shock renders the flow subsonic.
In the present case of flow past an angle on a finite solid surface, f we must
always take the former, i.e. the weak shock. It should be borne in mind that
this choice is really decided by the conditions of the flow far from the angle.
Fig. 95
In flow past a very acute angle (x small), the resulting shock wave must
obviously be very weak. It is natural to suppose that, as the angle increases,
the intensity of the shock increases monotonically ; this corresponds to a
movement along the arc QC of the shock polar (Fig. 50), from Q towards C.
We have also seen in §86 that the angle through which the velocity vector is
turned in a shock wave cannot exceed a certain value Xmax, which depends
on Mi. The flow pattern described above is therefore impossible if either of
the sides of the angle makes an angle greater than xmax with the direction of
the incident stream. In this case the gas flow near the angle must be sub
sonic; this is achieved by the appearance of a shock wave somewhere in front
of the angle (see §114). Since xmax increases monotonically with Mi, we can
also say that, for a given value of the angle x, Mi for the incident stream must
be greater than a certain value Mi min .
Finally it may be mentioned that, if the sides of the angle are situated,
relative to the incident stream, as shown in Fig. 96, then a shock wave is of
course formed on only one side of the angle ; the stream is turned on the other
side by a rarefaction wave.
PROBLEM
Determine the position and intensity of the shock wave in flow past a very small angle
(X <^ 1) for very large values of M : ( > 1/x).
t The purely formal problem of flow past a wedge formed by the intersection of two infinite planes
is of no physical interest.
418 The Intersection of Surf aces of Discontinuity §105
Solution. For x "^ 1 > the shock polar gives two values of <f>, one close to zero and the other
close to \tt. The weak shock which we require corresponds to the former value, which is
Kr+l)x; see §86, Problem 1. The ratio of pressures is, by (86.9), p^p x = JyCy+^MiV
The value of M behind the shock is
1
M 2 = 
X \J y(yl)'
i.e. it is still large compared with unity, but not large compared with 1/x
Fig. 96
§105. Flow past a conical obstacle
The problem of steady supersonic flow near a pointed projection on the
surface of a body is threedimensional, and is very much more complicated
than that of flow past an angle with a line vertex. No complete general in
vestigation of the former problem has yet been made. The only problem
that has been completely solved is that of axially symmetric flow past a pro
jecting point, and we shall discuss this case.
Near its vertex, an axially symmetric projection can be regarded as a right
cone of circular crosssection, and so the problem consists in investigating the
flow of a uniform stream past a cone whose axis is in the direction of inci
dence. The flow pattern is qualitatively as follows.
As in the analogous problem of flow past a twodimensional angle, a shock
wave must be formed (A. Busemann 1929), and it is evident from symmetry
that this shock is a conical surface coaxial with the cone and having the same
vertex (Fig. 97 shows the crosssection of the cone by a plane through its
axis). Unlike what happens in the twodimensional case, however, the shock
wave does not turn the gas velocity through the whole angle x necessary for
the gas to flow along the surface of the cone (2^ being the vertical angle of the
cone). After passing through the surface of discontinuity, the streamlines
are curved, and asymptotically approach the generators of the cone. This
curvature is accompanied by a continuous increase in density (besides the
increase which occurs at the shock itself) and by a corresponding decrease in
the velocity. Immediately behind the shock wave, the velocity is in general
still supersonic (as in the twodimensional case, it is determined by the
"supersonic" part of the shock polar), but on the surface of the cone it
§105
Flow past a conical obstacle
419
may become subsonic. As in the twodimensional case, for every value of
the Mach number Mi = vijc\ for the incident stream, there is a limiting value
Xmax for the angle of the cone, above which this type of flow becomes impos
sible.
The conical shock wave intersects all streamlines in the incident flow at
the same angle, and is therefore of constant intensity. Hence it follows (see
§106) that we have isentropic potential flow behind the shock wave also.
Fig. 97
From the symmetry of the problem and its similarity properties (there are
no characteristic constant lengths in the conditions imposed), it is evident
that the distribution of all quantities (velocity, pressure) in the flow behind
the shock wave will depend only on the angle 6 which the radius vector from
the vertex of the cone to the point considered makes with the axis of the cone
(the #axis in Fig. 97). Accordingly, the equations of motion are ordinary
differential equations; the boundary conditions on these equations at the
shock wave are determined by the equation of the shock polar, while those
at the surface of the cone are that the velocity should be parallel to the
generators. These equations, however, cannot be integrated analytically, and
have to be solved numerically. We refer the reader elsewheref for the results
of the calculations, and merely give the curve (Fig. 51, §86) which shows the
maximum possible angle xmax as a function of Mi. We may also mention
that, as Mi > 1, the angle xmax tends to zero:
Xmax = constantx V[(Mil)/(y+l)],
(105.1)
t For example, N. E. Kochin, I. A. Kibel* and N. V. Roze, Theoretical Hydromechanics (Teoretich
eskaya gidromekhanika), Part 2, 3rd ed., p. 193, Moscow 1948; L. Howarth ed., Modern Developments
in Fluid Dynamics: High Speed Flow, vol, 1, ch. 5, Oxford 1953.
420 The Intersection of Surfaces of Discontinuity §105
as may be deduced from the general law of transonic similarity (118.11);
the constant is independent both of Mi and of the gas involved.
An analytical solution of the problem of flow past a cone is possible only
in the limit of small vertical angles. It is evident that in this case the gas
velocity nowhere differs greatly from the velocity vi of the incident stream.
Denoting by v the small difference between the gas velocity at the point
considered and vi, and using its potential <f>, we can apply the linearised
equation (106.4); if we take cylindrical coordinates x, r, co with the polar
axis along the axis of the cone (co being the polar angle), this equation
becomes
or, for an axially symmetric solution,
1 8 / 86 \ 8 2 6
r 8r\ 8r / 8x 2
where
5= V(* 2 l) (105.4)
In order that the velocity distribution should be a function of 6 only, the
potential must be of the form 6 — x f{£)> where g = rjx = tan 0. Sub
stituting this, we obtain for the function /(£) the equation
£(l/*W"+/ f = 0,
of which the solution is elementary. The trivial solution / = constant
corresponds to a uniform flow; the other solution is
/ = constant x y(l  PH 2 )  cosh~i(l/^)].
The boundary condition on the surface of the cone (i.e. for f = tan x ~ x)
is
vrl{vi+v x ) « (l/©i)S0/0r = x (105.5)
or/ ' = vix Hence the constant is vix 2 , and we have the following expression
r the potential in the region x > fir :f
<f> = v lx *[y/{pP  £2 r 2) _ x co&hi(x/pr)]. (105.6)
It should be noticed that cf> has a logarithmic singularity for r + 0.
We can now find the velocity components :
v x = — vix 2 cosh 1 ( x I Br),
K (105.7)
v r = (vix 2 /r)\/(x 2 — 6 2 r 2 ).
The pressure on the surface of the cone is calculated from formula (106.5);
t In this approximation, the cone x = ]3r is a surface of weak discontinuity.
§105 Flow past a conical obstacle 421
since <j> has a logarithmic singularity for r > 0, the velocity v r on the surface
of the cone (i.e. for small r) is large compared with v x > and therefore we need
retain only the term in v r 2 in the formula for the pressure. The result is
PPi = />i^iV[log(2//?x)i]. (105.8)
All these formulae, which have been derived by means of a linearised theory,
cease to be valid for large Mi, comparable with 1/x (see §119).
The flow past a cone of arbitrary crosssection (the angle of attack being
not necessarily zero) is a similarity flow, like the symmetrical flow past a
circular cone. It has no characteristic length parameters, and so the velocity
distribution can be a function only of the ratios yjx, zjx of the coordinates,
i.e. it is constant along any straight line through the origin (the vertex of the
cone). Such similarity flows are called conical flows.f
PROBLEM
Determine the flow past a cone of small vertical angle 2x placed at a small angle of attack a
(C. Ferrari 1937)4
Solution. We take the axis of the cone (not the direction of the main stream) as the #axis ;
the linearised equation (105.2) for the potential is unchanged if higherorder quantities
('— < a<f>) are neglected, and the potential determines the gas velocity as V] +grad <f>. The boun
dary condition on the surface of the cone is
v\ sin a cos co + v r 1 8<f>
x a cos co \ & x
v\ cosa+^s vi 8r
We seek <f> as a sum :
<f> = <p> (x, r) + cos co • <P (x, r), (1)
where <£ (1 > is the expression (105.6), and <f> (2) satisfies the boundary condition 8<f> {i) l8r = — i^a.
The function <f>^ can be written as rf(r/x) and, substituting r/cos to in equation (105.2), we
obtain for / the equation
f/"(^ 2 l)+/'(2^ 2 3) = 0.
The trivial solution / = constant corresponds to a uniform stream incident (with velocity
i> x a) in a direction perpendicular to the axis of the cone ; the other solution leads to
<P = vtfxMWfrWix 2  P 2 r 2 ) fir coshi(xlpr)].
The gas velocity is Vj+v'^+v* 2 ), where v (2 > = grad <f>W and v (1) is given by formulae
(105 .7). The pressure is calculated from the formula
p —pi = — Jpi{(^i cos a + 8<f>/dx) 2 +
+ (vi sin a cos co + 8(f>Jdr) 2 + ( — v\ sin a sin co + 8</>/r 8co) 2 — v\ 2 }
in whi ch the secondorder terms in a and x must be retained. The pressure on the surface
of the cone is found to be given by
PPi = PiVi*{x 2 \og{2lp x )\{ x 2 + «. 2 )
— lax cos w + a 2 cos 2co).
f A detailed account of various problems concerning these flows is given by E. Carafoli, High
Speed Aerodynamics {Compressible Flow), Pergamon Press, London 1958.
J The solution of the same problem for any thin solid of revolution is given by F. I. Frankl'
and E. A. Karpovich, Gas Dynamics of Thin Bodies, §27, Interscience, New York 1953.
CHAPTER XII
TWODIMENSIONAL GAS FLOW
§106. Potential flow of a gas
In what follows we shall meet with many important cases where the flow of
a gas can be regarded as potential flow almost everywhere. Here we shall
derive the general equations of potential flow and discuss the question of their
validity.
After passing through a shock wave, potential flow of a gas usually becomes
rotational flow. An exception, however, is formed by cases where a steady
potential flow passes through a shock wave whose intensity is constant over
its area; such, for example, is the case where a uniform stream passes through
a shock wave intersecting every streamline at the same angle, f The flow
behind the shock wave is then potential flow also. To prove this, we use
Euler's equation in the form
grada 2 vxcurlv = (1/p) grad/>
(cf. (2.10)), or
grad(w + %v 2 ) vxcurlv= T grad s,
where we have used the thermodynamic identity dzo = Tds + dp/p. In
potential flow, however, w + %v 2 = constant in front of the shock wave, and
this quantity is continuous at the shock; it is therefore constant everywhere
behind the shock wave, so that
vxcurlv = Tgrads. (106.1)
The potential flow in front of the shock wave is isentropic. In the general
case of an arbitrary shock wave, for which the discontinuity of entropy varies
over its surface, grad s # in the region behind the shock, and curl v is
therefore also not zero. If, however, the shock wave is of constant intensity,
then the discontinuity of entropy in it is constant, so that the flow behind the
shock is also isentropic, i.e. grad s = 0. From this it follows that either
curl v = or the vectors v and curl v are everywhere parallel. The latter,
however, is impossible; at the shock wave, v always has a nonzero normal
component, but the normal component of curl v is always zero (since it is
given by the tangential derivatives of the tangential velocity components,
which are continuous).
Another important case where potential flow continues despite the shock
wave is that of a weak shock. We have seen (§83) that in such a shock wave
f We have already met with this situation in connection with supersonic flow past a wedge or
cone (§§104, 105).
422
§106 Potential flow of a gas 423
the discontinuity of entropy is of the third order relative to the discontinuity
of pressure or velocity. We therefore see from (106.1) that curl v behind the
shock is also of the third order. This enables us to assume that we have
potential flow behind the shock wave, the error being of a high order of small
ness.
We shall now derive the general equation for the velocity potential in an
arbitrary steady potential flow of a gas. To do so, we eliminate the density
from the equation of continuity div(/>v) = p div v+ vgrad p = 0, using Euler's
equation
(vgrad)v = (l/p)grad/> = (c 2 //>)grad/>
and obtaining
c 2 divv— v«(v«grad)v = 0.
Introducing the velocity potential by v = grad <f> and expanding in components,
we obtain the equation
{c 2 ^)<f>x X + {c 2 ^)<j> y y+{c 2 ^)4> zz 
f lUo.Z)
— 2(<f> X <f>y(l> X y+<f>y<f> Z ff>y z +<f) z <l) X <f) zx ) — 0,
where the suffixes here denote partial derivatives. In particular, for two
dimensional flow we have
(c 2  <f> X 2 )<f>XX + (c 2  <f>y 2 )cf>yy ~ 2cf>x<f>y<f>xy = 0. (106.3)
In these equations, the velocity of sound must itself be expressed in terms of
the velocity ; this can in principle be done by means of Bernoulli's equation,
to + \v 2 = constant, and the isentropic equation, s = constant. For a perfect
gas, c as a function of v is given by formula (80.18).
Equation (106.2) is much simplified if the gas velocity nowhere differs
greatly in magnitude or direction from that of the stream incident from
infinity.) This implies that the shock waves (if any) are weak, and so the
potential flow is not destroyed.
As in similar cases previously, we denote by v the small difference between
the gas velocity at a given point and that of the main stream. Denoting the
latter by vi, we therefore write the total velocity as vi + v. The potential
<f> is taken to mean that of the velocity v: v = grad <f>. The equation for this
potential is obtained from (106.2) by substituting (f> xfy + xvi; we take the
#axis in the direction of the vector vi. We then regard ^ as a small quantity,
and omit all terms of order higher than the first, obtaining the following
linear equation:
8 2 6 8 2 <t> 8 2 J>
where Mi = vijci; the velocity of sound is, of course, given its value at
infinity.
t One such case was discussed in §105 (flow past a narrow cone), and others will be found in con
nection with gas flow past arbitrary thin bodies.
424 Twodimensional Gas Flow §106
The pressure at any point is determined in terms of the velocity in the
same approximation, by a formula which can be obtained as follows. We
regard p as a function of w (for given s), and use the fact that {dw\dp) s = 1/p,
writing p — pi « (dpldw) s (w — wi) = pi(ww{). From Bernoulli's equation
we have
wwi= _i[( Vl +v) 2 vi 2 ] « KV+^ 2 )^i%,
so that
ppi = pivivxipi(v y 2 +v z 2 ). (106.5)
In this expression the term in the squared transverse velocity must in general
be retained, since, in the region near the #axis (and, in particular, on the
surface of the body itself), the derivatives 8<f>Jdy y dtfrjdz may be large com
pared with d<j>Jdx.
Equation (106.4), however, is not valid if the number Mi is very close to
unity {transonic flow), so that the coefficient of the first term is small. It is
clear that, in this case, terms of higher order in the ^derivatives of <£ must
be retained. To derive the corresponding equation, we return to the original
equation (106.2); when the terms which are certainly small are neglected,
this becomes
*<j>xx + <t>yy + 4>zz = 0. (106.6)
(S)
In the present case, the velocity v x = v, and the velocity of sound c is
close to the critical velocity c* (v now denoting the total velocity). We can
therefore put cc* = (vc*) {dcjdv) v=Cif , or cv = {c*v)[\(dcjdv) v = c ^.
Using the fact that, for v = c = c^, we have by (80.4) dpjdv = —p]c, we
put (for v = c*)
dc dc dp p dc
dv dp dv c dp
so that
cv = [(c*v)/c]d(pc)/dp = a*(c*z>). (106.7)
We have here used the expression (92.9) for the derivative d(pc)jdp, while a*
denotes the value of a (95.2) for v = c*; for a perfect gas, a is constant, so
that a* = a = l(y+ 1). To the same accuracy, this equation can be written
as
v/c1 = a*(©/c*l). (106.8)
This gives the general relation between the Mach numbers M and M % in
transonic flow.
Using this formula, we can put
^ 2 ~ v2
2(1^*2^(1;)
§107 Steady simple waves 425
Finally, we introduce a new potential by the substitution <j> +Cx(x + <j>), so
that
^ = ^_l, *_f* *_* (106.9)
&c r* dy c* dz c*
Substituting these formulae in (106.6), we obtain the following final equation
for the velocity potential in a transonic flow (with the velocity everywhere
almost parallel to the #axis) :
dd> cN> 8U dU
2a*— L = L + L. (106.10)
8x dx 2 dy 2 dz 2
The properties of the gas appear here only through the constant a*. We shall
see later that this constant governs the entire dependence of the properties of
transonic flow on the nature of the gas.
The linearised equation (106.4) becomes invalid also in another limiting
case, that of very large values of Mi : however, the appearance of strong shock
waves has the result that potential flow cannot actually occur for such
values of Mi (see §119).
§107. Steady simple waves
Let us determine the general form of those solutions of the equations of
steady twodimensional supersonic gas flow which describe flows in which
there is a uniform planeparallel stream at infinity, which then turns through
an angle as it flows round a curved profile. We have already met a particular
case of such a solution in discussing the flow near an angle ; the flow considered
was essentially a planeparallel one along one side of the angle, which turned
at the vertex of the angle. In this particular solution all quantities (the two
velocity components, the pressure and the density) were functions of only
one variable, the angle <f>. Each of these quantities could therefore be expres
sed as a function of any other. Since this solution must be a particular case
of the required general solution, it is natural to seek the latter on the assump
tion that each of the quantities p, p, v x , v y (the plane of the motion being
taken as the ryplane) can be expressed as a function of any other. This
assumption is, of course, a very considerable restriction on the solution of
the equations of motion, and the solution thus obtained is not the general
integral of those equations. In the general case, each of the quantities p,
P* Vx> Vy> which are functions of the two coordinates x, y, can be expressed
as a function of any two of them.
Since we have a uniform stream at infinity, in which all quantities, and in
particular the entropy s, are constants, and since in steady flow of an ideal
fluid the entropy is constant along the streamlines, it is clear that s = constant
in all space if there are no shock waves in the gas, as we shall assume.
426 Twodimensional Gas Flow §107
Euler's equations and the equation of continuity are
dv x dv x 1 dp dv y 8v v 1 dp
v x — \ Vy—— = , v x h v v = ;
dx dy pdx dx V dy p dy'
d d
rip®*) + —fay) = 0.
dx dy
Writing the partial derivatives as Jacobians, we can convert these equations
to the form
8 ( v x,y) 8(v x ,x) 1 d(p,y)
V X Vy = ,
d{x,y) d(x,y) p d(x,y)
d(v yy y) d(vy,x) 1 8(p,x)
v x Vy = ;
d(x,y) d{x,y) p d(x,y)
d(pv x ,y) d{pv yi x)
8(x,y) 8(x,y)
We now take (say) x and p as independent variables. In order to effect this
transformation, we need only multiply the above equations by d(x, y)jd{x, p),
obtaining the same equations except that 8(x,p) replaces d(x,y) in the
denominator of each Jacobian. We now expand the Jacobians, bearing in
mind that all the quantities />, v x , v y are assumed to be functions of p but not
of x, so that their partial derivatives with respect to x are zero. We then
obtain
/ ty\ dv x 1 dy / dy\ dvy 1
r*^/ # =  P Vx \ Vy  v *T x ) & =  ?
\ dx) dp ^\ dp dx dp)
Here dy\dx denotes (dyjdx) p . All the quantities in these equations except
dyjdx are functions of p only, by hypothesis, and x does not appear explicitly.
We can therefore conclude, first of all, that dyjdx also is a function of p only:
(dyjdx) p = fi{p), whence
y = xfi(p)+f2(p), (107.1)
where /2(/>) is an arbitrary function of the pressure.
No further calculations are necessary if we use the particular solution,
already known, for a rarefaction wave in flow past an angle (§§101, 104).
It will be recalled that, in this solution, all quantities (including the pressure)
are constants along any straight line (characteristic) through the vertex of
the angle. This particular solution evidently corresponds to the case where the
arbitrary function f2(p) in the general expression (107.1) is identically zero.
The function f±(p) is determined by the formulae obtained in §101.
§107
Steady simple waves
427
Equation (107.1) for various constant/) gives a family of straight lines in
the ryplane. These lines intersect the streamlines at every point at the Mach
angle. This is seen immediately from the fact that the lines y = xfi(p) in
the particular solution with/ 2 = have this property. Thus one of the fami
lies of characteristics (those leaving the surface of the body) consists, in the
general case, of straight lines along which all quantities remain constant;
these lines, however, are no longer concurrent.
The properties of the flow described above are, mathematically, entirely
analogous to those of onedimensional simple waves, in which one family of
characteristics is a family of straight lines in the xt plane (see §§94, 96, 97).
'////V////////V//V/V////
Fig. 98
Hence the class of flows under consideration occupies the same place in the
theory of steady (supersonic) twodimensional flow as do simple waves in
nonsteady onedimensional flow. On account of this analogy, such flows
are also called simple waves; in particular, the rarefaction wave which cor
responds to the case/2 = is called a centred simple wave.
As in the nonsteady case, one of the most important properties of steady
simple waves is that the flow in any region of the #yplane bounded by a
region of uniform flow is a simple wave (cf. §97).
We shall now show how the simple wave corresponding to flow round a
given profile can be constructed. Fig. 98 shows the profile in question;
to the left of the point O it is straight, but to the right it begins to curve. In
supersonic flow the effect of the curvature is, of course, propagated only
downstream of the characteristic OA which leaves the point O. Hence the flow
to the left of this characteristic is uniform; we denote by the suffix 1 quanti
ties pertaining to this region. All the characteristics there are parallel and at
an angle to the *axis which is equal to the Mach angle a x = sin 1 (ci/*;i).
In formulae (101.12)(101.15), the angle <f> of the characteristics is measured
from the line on which v = c = C*. This means (cf. §104) that the charac
teristic OA must have a value of <£ given by
x /^ +1 i Cl
h = / — rcos 1— ,
428 Twodimensional Gas Flow §107
and the angle <f> is to be measured from OA' (Fig. 98). The angle between
the characteristics and the #axis is then 0*^, where <f>* = ai + ^x. Accord
ing to formulae (101.12)(101.15), the velocity and pressure are given in
terms of (j> by
v x = v cos 9, v y = v sin0, (107.2)
* 2 = '4 1 + ^ sin V^]' (107.3)
6 = ^^tani(ygcoty^), (107.4)
p = p* cos2r/(rD j^—6. (107.5)
V y+ 1
The equation of the characteristics can be written
y = xtan((f>*cf>) + F(cf>). (107.6)
The arbitrary function F(<f>) is determined as follows when the form of the
profile is given. Let the latter be Y = Y(X), where X and Y are the co
ordinates of points on it. At the surface, the gas velocity is tangential, i.e.
tan0 = dY/dX. (107.7)
The equation of the line through the point (X, Y) at an angle <f> m <f> to the
xaxis is
yY = {xX) tan(<£ # <£).
This equation is the same as (107.6) if we put
F(<f>) = YXtanO^^). (107.8)
Starting from the given equation Y = Y(X) and equation (107.7), we express
the form of the profile in parametric equations X = X(6), Y = Y(6), the
parameter being the inclination 6 of the tangent. Substituting 6 in terms of <f>
from (107.4), we obtain X and Fas functions of <f>; finally, substituting these
in (107.8), we obtain the required function F((f>).
In flow past a convex surface, the angle 6 between the velocity vector
and the xaxis decreases downstream (Fig. 98), and the angle <£*0 between
the characteristic and the xaxis therefore decreases monotonically also (we
always mean the characteristic leaving the surface). For this reason, the
characteristics do not intersect (in the region of flow, that is). Thus, in the
region downstream of the characteristic OA (which is a weak discontinuity),
we have a continuous (no shock waves) and increasingly rarefied flow.
The situation is different in flow past a concave profile. Here the inclina
tion 6 of the tangent increases downstream, and therefore so does the inclina
tion of the characteristics. Consequently, the characteristics intersect in the
§107
Steady simple waves
429
region of flow. On different nonparallel characteristics, however, all
quantities (velocity, pressure, etc.) have different values. Thus all these
quantities become manyvalued at points where characteristics intersect,
which is physically impossible. We have already met a similar phenomenon
in connection with a nonsteady onedimensional simple compression wave
(§94). As in that case, it signifies that in reality a shock wave is formed.
The position of the discontinuity cannot be completely determined from the
solution under consideration, since this was derived on the assumption that
there are no discontinuities. The only result that can be obtained is the
place where the shock wave begins (the point O in Fig. 99, where the shock
is shown by the continuous line OB). It is the point of intersection of charac
teristics whose streamline lies nearest to the surface of the body. On stream
lines passing below O (i.e. nearer to the surface) the solution is everywhere
onevalued; its manyvaluedness "begins" at O. The equations for the co
ordinates #o, yo of this point can be obtained in the same way as the cor
responding equations which determine the time and place of formation of
the discontinuity in a onedimensional nonsteady simple wave. If we regard
the inclination of the characteristics of a function of the coordinates (x, y)
of points through which they pass, then this function becomes manyvalued
when x and y exceed certain values xq, yo. In §94 the situation was the same
in relation to the function v(x, t), and so we need not repeat the arguments
used there, but can write down immediately the equations
{py\H)x = 0, {&y\m* = 0, (107.9)
which now determine the place of formation of the shock wave. Mathemati
cally, this point is a cusp on the envelope of the family of straight charac
teristics (cf. §96).
In flow past a concave profile, the simple wave exists along streamlines
passing above O as far as the points where these lines intersect the shock
wave. The streamlines passing below O do not intersect the shock wave at
all, but we cannot conclude from this that the solution in question is valid
at all points on these streamlines. The reason is that the shock wave has a
430 Twodimensional Gas Flow §108
perturbing effect even on the gas which flows along these streamlines, and so
alters the flow from what it would be in the absence of the shock wave. By
a property of supersonic flow, however, these perturbations reach only the
gas downstream of the characteristic OA (of the second family) which leaves
the point where the shock wave begins. Thus the solution under considera
tion is valid everywhere to the left of AOB. The line OA itself is a weak
discontinuity. We see that there cannot be a continuous (no shock waves)
simple compression wave everywhere in flow past a concave surface, which
would correspond to the simple rarefaction wave in flow past a convex surface.
The shock wave formed in flow past a concave profile is an example of a
shock which "begins" at a point inside the stream, away from the solid walls.
The point where the shock begins has some general properties, which may be
noted here. At the point itself the intensity of the shock wave is zero, and
near the point it is small. In a weak shock wave, however, the discontinuities
of entropy and vorticity are of the third order of smallness, and so the change
in the flow on passing through the shock differs from a continuous potential
isentropic change only by quantities of the third order. Hence it follows
that, in the weak discontinuities which leave the point where the shock wave
begins, only the third derivatives of the various quantities can be discontinu
ous. There will in general be two such discontinuities : a weak discontinuity
coinciding with the characteristic, and a weak tangential discontinuity coin
ciding with the streamline (see the end of §89).
§108. Chaplygin's equation: the general problem of steady two
dimensional gas flow
Having dealt with steady simple waves, let us now consider the general
problem of an arbitrary steady plane potential flow. We assume that the flow
is isentropic and contains no shock waves.
As was first shown by S. A. Chaplygin in 1902, it is possible to reduce
this problem to the solution of a single linear partial differential equation.
This is achieved by means of a transformation to new independent variables,
the velocity components v x , v y ; this transformation is often called the hodo
graph transformation, the ^%plane being called the hodograph plane and the
xyplane the physical plane.
For potential flow we can replace Euler's equations by their first integral,
Bernoulli's equation:
w+&2 = w . (108.1)
The equation of continuity is
T&*) + TV*) = 0. (108.2)
dx By
For the differential of the velocity potential <f> we have d<j> = v x dx+v y dy.
We transform from the independent variables x, y to the new variables v x , v y
§108 Chaplygiris equation 431
by Legendre's transformation, putting
d<f> = d(xv x )xdv x + d(yvy)ydv y ,
introducing the function
d> = cff + xvz+yvy, (108.3)
and obtaining
d<I> = xdv x +ydvy,
where <1> is regarded as a function of v x and v y . Hence
x = d<bldv x , y = dQjdvy. (108.4)
It is more convenient, however, to use, not the Cartesian components of the
velocity, but its magnitude v and the angle 9 which it makes with the a:axis :
•v x = vcosd, Vy = vsin6. (108.5)
The appropriate transformation of the derivatives gives, instead of (108.4),
3<J> sin0 8® . S<D cos0 dd> /ino , x
x = cos e — , y = sin 0—  + — . (108.6)
8v v dQ 8v v 86
The relation between the potential <p and the function O is given by the
simple formula
<£ = ®+vd®ldv. (108.7)
Finally, in order to obtain the equation which determines the function
0(a, 0), we must transform the equation of continuity (108.2) to the new
variables. Writing the derivatives as Jacob ians:
d{pwx,y) d(pvy,x) _
8(x,y) d(x,y)
multiplying by d(x } y)]d(v, 6) and substituting (108.5), we have
8(pv cos 0, y) 8{pv sin 6, x) _
8(v,6) d(v, 0)
To expand these Jacobians, we must substitute (108.6) for x and y. Further
more, since the entropy s is a given constant, if we express the density as a
function of s and zv and substitute to = Wq  %v 2 we find that the density can
be written as a function of v only: p = p(v). We therefore obtain, after a
simple calculation, the equation
d(pv) I 8® 1 S2<X> \ 32$
JlL { + + p v — — = 0.
dv \8v v 86 2 J dv 2
According to (80.5),
d(pv) / v* \
432 Twodimensional Gas Flow §108
and so we have finally Chaplygin's equation for the function <!>(#, 6) :
6 2 <J> v 2 8 2 <& 3<X>
+ 7^5 ^ + ^^T = ° (108.8)
£02 1^2 &,2 ^
Here the velocity of sound is a known function c(v), determined by the
equation of state of the gas together with Bernoulli's equation.
The equation (108.8), together with the relations (108.6), is equivalent to
the equations of motion. Thus the problem of solving the nonlinear equa
tions of motion is reduced to the solution of a linear equation for the function
®(v, 6). It is true that the boundary conditions on this equation are non
linear. These conditions are as follows. At the surface of the body, the gas
velocity must be tangential. Expressing the equation of the surface in the
parametric form X = X(6), Y = Y(0) (as in §107), and substituting X and
Y in place of x and y in (108.6), we obtain two equations, which must be
satisfied for all values of 0; this is not possible for every function ®(v, 6).
The boundary condition is, in fact, that these two equations are compatible
for all 6, i.e. one of them must be deducible from the other.
The satisfying of the boundary conditions, however, does not ensure that
the resulting solution of Chaplygin's equation determines a flow that is
actually possible everywhere in the physical plane. The following condition
must also be met: the Jacobian A = 8(x,y)jd(d, v) must nowhere be zero,
except in the trivial case when all its four component derivatives vanish.
It is easy to see that, unless this condition holds, the solution becomes com
plex when we pass through the line (called the limiting line) in the xyplane
given by the equation A = O.f For, let A = on the line v = v (6), and
suppose that (dyjdd) v ^ 0. Then we have
a(\  8(x,y) 8{v,d)  d ^ y) = ( 8x \ =0
\8yJ v 8(v,d) d(v,y) 8{v,y) \8v! y
Hence we see that, near the limiting line, v is determined as a function of x
(for given y) by
x — xq — %(d 2 xldv 2 ) y (v — vo) 2 ,
and v becomes complex on one side or the other of the limiting line. J
It is easy to see that a limiting line can occur only in regions of supersonic
flow. A direct calculation, using the relations (108.6) and equation (108.8),
gives
1
A = 
v
e 2 d> l ao>\ 2 v 2 i »
+
86 dv v 89 J \v 2 \c 2 \ 8v 2
(108.9)
t There is no objection to a passage through points where A becomes infinite. If 1/A = on some
line, this merely means that the correspondence between the xy and vQ planes is no longer oneto
one : in going round the xyplane, we cover some part of the w#plane two or three times.
X This result clearly remains valid even if (8 2 xl8v 2 ) y vanishes with A but {dxjdv) y again changes
sign for v = v , i.e. the difference x—x is proportional to a higher even power of v — vq.
§109 Characteristics in steady twodimensional flow 433
It is clear that, for v ^ c, A > 0, and A can become zero only if v > c.
The appearance of limiting lines in the solution of Chaplygin's equation
indicates that, under the given conditions, a continuous flow throughout the
region is impossible, and shock waves must occur. It should be emphasised,
however, that the position of these shocks is not the same as that of the
limiting lines.
In §107 we discussed the particular case of steady twodimensional super
sonic flow (a simple wave), which is characterised by the fact that the velocity
in it is a function only of its direction: v — v(6). This solution cannot be
obtained from Chaplygin's equation, since 1/A = 0, and the solution is
"lost" when the equation of continuity is multiplied by the Jacob ian A in the
transformation to the hodograph plane. The situation is exactly analogous
to that found in the theory of nonsteady onedimensional flow. The re
marks made in §98 concerning the relation between the simple wave and the
general integral of equation (98.2) are wholly applicable to the relation be
tween the steady simple wave and the general integral of Chaplygin's equation.
The fact that the Jacobian A is positive in subsonic flow enables us to
demonstrate an interesting theorem due to A. A. Nikol'skii and G. I.
Taganov (1946). We have identically
1 _ 8(0, v) _ 8(6, v) 8(x,v)
A 8(x,y) 8(x,v) 8(x,y)
or
A \8x/ v \dyj
In a subsonic flow A > 0, and we see that the derivatives (ddjdx) v and
(dvjdy) x have the same sign. This has a simple geometrical significance : if
we move along a line v = constant = vq, with the region v < vo to the right,
the angle 6 increases monotonically, i.e. the velocity vector turns always
counterclockwise. This result holds, in particular, for the line of transition
between subsonic and supersonic flow, on which v = c = c%.
In conclusion, we may give Chaplygin's equation for a perfect gas, writing
c explicitly in terms of v :
Q2® l (y l)„2 /(y+1K 2 £2$ g<J>
+ v 2 V v = 0. (108.11)
_ =   ( _ . (108.10)
86 2 \v 2 \c* 2 8v 2 8v
This equation has a family of particular integrals expressible in terms of
hypergeometric functions.]"
§109. Characteristics in steady twodimensional flow
Some general properties of characteristics in steady (supersonic) two
dimensional flow have already been discussed in §79. We shall now derive
t See, for instance, L. I. Sedov, Twodimensional Problems of Hydrodynamics and Aerodynamics
(Ploskie zadachi gidrodinamiki i aerodinamiki), Moscow 1950.
434 Twodimensional Gas Flow §109
the equations which give the characteristics in terms of a given solution of the
equations of motion.
In steady twodimensional supersonic flow there are, in general, three
families of characteristics. All small disturbances, except those of entropy
and vorticity, are propagated along two of these families (which we call the
characteristics C + and C_); disturbances of entropy and vorticity are pro
pagated along characteristics (C ) of the third family, which coincide with
the streamlines. For a given flow, the streamlines are known, and the problem
is to determine the characteristics belonging to the first two families.
The directions of the characteristics C + and C_ passing through each point
in the plane lie on opposite sides of the streamline through that point, and make
with it an angle equal to the local value of the Mach angle a (Fig. 41, §79).
We denote by m the slope of the streamline at a given point, and by m +> m
the slopes of the characteristics C + , C_. Then we have
whence
m + — mo
1 + motn +
m—niQ
= tana, — =
1 + mom
mo + tan a
1 + mo tan a
— tan a,
the upper signs everywhere relate to C + and the lower to C_. Substituting
m = Vyjv x , tana = cj^{v 2 c 2 ) and simplifying, we obtain the following
expression for the slopes of the characteristics :
/ dy\ v x v y ± c\/(v 2  c 2 )
If the velocity distribution is known, this is a differential equation which
determines the characteristics C + and C.f
Besides the characteristics in the xyplane, we may consider those in the
hodograph plane, which are especially useful in the discussion of isentropic
potential flow; we shall take this case in what follows. Mathematically, these
are the characteristics of Chaplygin's equation (108.8), which is of hyperbolic
type for v > c. Following the general method familiar in mathematical
physics (see §96), we form from the coefficients the equation of the charac
teristics :
dv* + dd 2 v 2 /(lv 2 lc*) = 0,
or
/d0\ 1 \iv 2 \
Ur^Jh 1 ) (m2)
t Equation (109.1) also determines the characteristics for steady axially symmetric flow if v„ and
y are replaced by v T and r, where r is the cylindrical coordinate (the distance from the axis
of symmetry, which is the *axis); it is clear that the derivation is unchanged if we consider, instead of
the acyplane, an xrplane through the axis of symmetry.
§109
Characteristics in steady twodimensional flow
435
The characteristics given by this equation do not depend on the particular
solution of Chaplygin's equation considered, because the coefficients in that
equation are independent of O. The characteristics in the hodograph plane
are a transformation of the characteristics C + and C in the physical plane,
and we call them respectively the characteristics V + and T_, in accordance
with the signs in (109.2).
The integration of equation (109.2) gives relations of the form J + (v, d)
= constant, J(v, 6) = constant. The functions / + and /_ are quantities
which remain constant along the characteristics C + and C (i.e. Riemann
invariants). For a perfect gas, equation (109.2) can be integrated explicitly.
Fig. 100
There is, however, no need to go through the calculations, since the result
can be seen from formulae (107.3) and (107.4). For, according to the general
properties of simple waves (see §97), the dependence of v on 6 for a simple
wave is given